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1607

1607.02567_arXiv.txt

We investigate the running cosmological constant model with dark energy linearly proportional to the Hubble parameter, $\Lambda = \sigma H + \Lambda_0$, in which the $\Lambda$CDM limit is recovered by taking $\sigma=0$. We derive the linear perturbation equations of gravity under the Friedmann-Lema\"itre-Robertson-Walker cosmology, and show the power spectra of the CMB temperature and matter density distribution. By using the Markov chain Monte Carlo method, we fit the model to the current observational data and find that $\sigma H_0/ \Lambda_0 \lesssim 2.63 \times 10^{-2}$ and $6.74 \times 10^{-2}$ for $\Lambda(t)$ coupled to matter and radiation-matter, respectively, along with constraints on other cosmological parameters.

\label{sec:introduction} The type-Ia supernova observations~\cite{Riess:1998cb,Perlmutter:1998np} have shown that our universe is undergoing a late-time accelerating expansion, which is caused by Dark Energy~\cite{Copeland:2006wr}. The simplest way to realize such a late-time accelerating mechanism is to introduce a cosmological constant to the gravitational theory, such as that in the $\Lambda$CDM model. This model fits current cosmological observations very well, but exists several difficulties, such as the ``fine-tuning"~\cite{Review1, WBook} and ``coincidence''~\cite{CC} problems. In this work, we will concentrate on the latter problem~\cite{CP}, which has been extensively explored in the literature. One of the popular attempts is the running $\Lambda$ model, in which the cosmological constant evolves in time and decays to matter in the evolution of the universe~\cite{Ozer:1985ws, Carvalho:1991ut, Lima:1994gi, Lima:1995ea, Overduin:1998zv, Carneiro:2004iz, Shapiro:2009dh, Geng:2016epb, Bauer:2005rpa, Shapiro:2004ch, Dymnikova:2001ga, Alcaniz:2005dg, Barrow:2006hia}, so that the present energy densities of dark energy and dark matter are of the same order of magnitude. Its observational applications have been investigated in Ref.~\cite{EspanaBonet:2003vk, Tamayo:2015qla, Sola:2016vis}. In our study, we are interested in the specific model with $\Lambda = \sigma H$~\cite{Borges:2005qs, Borges:2007bh, Carneiro:2007bf, Borges:2008ii, Zimdahl:2011ae, Alcaniz:2012mh}, which would originate from the theory with the QCD vacuum condensation associated with the chiral phase transition~\cite{Schutzhold:2002pr, Klinkhamer:2009nn, Banerjee:2003fg, Ohta:2010in, Cai:2010uf}. In this scenario, the cosmological constant decays to matter (non-relativistic) and radiation (relativistic), leading to a large number of particles created in the cosmological evolution. Without loss of generality, we phenomenologically extend this model to include that $\Lambda$ additionally couples to radiation with $\Lambda = \sigma H + \Lambda_0$~\cite{Basilakos:2009wi, Costa:2012xw, Gomez-Valent:2014rxa}, in which the $\Lambda$CDM limit can be realized if $\sigma=0$. In this scenario, when dark energy dominates the universe, the decay rate of $\Lambda$ is reduced, and the late-time accelerating phase occurs, describing perfectly the evolution history of the universe. As a result, it is reasonable to go further to analyze the cosmological behavior of this model at the sub-horizon scale. In this paper, we examine the matter power spectrum $P(k)$ and CMB temperature perturbations in the linear perturbation theory of gravity. By using the Markov chain Monte Carlo (MCMC) method, we perform the global fit from the current observational data and constrain the model. This paper is organized as follows: In Sec.~\ref{sec:model}, we introduce the $\Lambda(t)$CDM model and review its background cosmological evolutions. In Sec.~\ref{sec:perturbation}, we calculate the linear perturbation theory and illustrate the power spectra of the matter distribution and CMB temperature by the {\bf CAMB} program~\cite{Lewis:1999bs}. In Sec.~\ref{sec:constraints}, we use the {\bf CosmoMC} package~\cite{Lewis:2002ah} to fit the model from the observational data and show the constraints on cosmological parameters. Our conclusions are presented in Sec.~\ref{sec:conclusion}.

\label{sec:conclusion} We have investigated the $\Lambda(t)$CDM model with the dark energy decaying to both matter and radiation, in which $\Lambda(t)= \sigma H + \Lambda_0$. Although this scenario is suitable to describe the late-time accelerating universe at the background level, the linear perturbation analyses of the matter power and CMB temperature spectra have set a strong constraint on the model parameter $\lambda_1$ in Eq.~(\ref{eq:rhol}). Explicitly, by performing the global fit from the observational data, we have obtained that $\lambda_1 \simeq \sigma H_0/\Lambda_0 \lesssim 6.68 \times 10^{-2}$ ($2.63 \times 10^{-2}$) and $\chi^2_{\Lambda(t)\mathrm{CDM}}=13546.5 (13545.7) \gtrsim \chi^2_{\Lambda\mathrm{CDM}}=13545.3$ for $C_r = 1 (0)$, implying that the current data prefers the $\Lambda$CDM limit. Constraints on other cosmological parameters in both $\Lambda(t)$CDM and $\Lambda$CDM models have been also given in Table~\ref{tab:2}.

1607.02567

1607

1607.00612_arXiv.txt

name}{Aiiioaoey} \begin{document} \title{Dark Matter Density Spikes around Primordial Black Holes} \author{Yu. N. Eroshenko}\thanks{e-mail: eroshenko@inr.ac.ru} \affiliation{Institute for Nuclear Research, Russian Academy of Sciences, pr. 60-letiya Oktyabrya 7a, Moscow, 117312 Russia} \date{\today} \begin{abstract} We show that density spikes begin to form from dark matter particles around primordial black holes immediately after their formation at the radiation-dominated cosmological stage. This follows from the fact that in the thermal velocity distribution of particles there are particles with low velocities that remain in finite orbits around black holes and are not involved in the cosmological expansion. The accumulation of such particles near black holes gives rise to density spikes. These spikes are considerably denser than those that are formed later by the mechanism of secondary accretion. The density spikes must be bright gamma-ray sources. Comparison of the calculated signal from particle annihilation with the Fermi-LAT data constrains the present-day cosmological density parameter for primordial black holes with masses $M_{\rm BH}\geq10^{-8}M_\odot$ from above by values from $\Omega_{\rm BH}\leq1$ to $\Omega_{\rm BH}\leq10^{-8}$, depending on $M_{\rm BH}$. These constraints are several orders of magnitude more stringent than other known constraints.

Primordial black holes (PBHs), the possibility of whose formation was predicted in \cite{p41} and \cite{p22}, can give valuable information about processes in the early Universe \cite{p12,p14,p3}, in particular, about the shape of the perturbation spectrum on small scales \cite{p24}. The quantum evaporation of low-mass PBHs is important from the viewpoint of investigating fundamental processes at high energies \cite{p9} and can have significance for the theory of primordial nucleosynthesis and gamma-ray astronomy. In addition, PBHs can offer new possibilities for the formation of quasars at high $z$ \cite{p16} and for baryonic objects with chemical peculiarities\cite{p18,p19}. Being captured by neutron stars, PBHs can affect their evolution, which gives a constraint on the number of PBHs \cite{p10}. In this paper, we will discuss only the PBHs that are formed during the collapses of adiabatic density perturbations, when a mixture of relativistic particles collapses into a PBH at the instant the perturbation crosses the cosmological horizon \cite{p11}. Note, however, that other PBH formation models have also been proposed at early dust-like stages \cite{p25,p40} or through the collapses of domain walls \cite{p4}, \cite{p26,p37}. PBHs can themselves represent dark matter (DM) \cite{p23} if they are formed in sufficiently large quantities, but they can also serve as seeds for the formation of DM clumps \cite{p15,p35,p30,p36,p28,p38}. Secondary accretion (generally, this mechanism was developed in cold DM onto a PBH \cite{p7}, when DM flows toward the PBH and is virialized at some radius to form a halo, is usually considered in investigating DM clumps around PBHs. In this paper, we will show that the DM density around PBHs can reach much greater values than that under secondary accretion. This stems from the fact that in the thermal velocity distribution there are DM particles with low velocities that remain in finite orbits around PBHs and are not involved in the overall cosmological expansion. The accumulation of such particles around PBHs gives rise to density spikes (halos). Two regimes of density spike formation around PBHs are possible at the radiation-dominated stage. In the first case, which occurs for PBHs with masses $M_{\rm BH}\leq40M_\odot$, PBHs are formed before the kinetic decoupling of DM particles (under the assumption that the DM particles are neutralinos with masses $m\sim70$~GeV). In the interval between the PBH formation and kinetic decoupling, a DM overdensity has time to be formed around the PBH. As will be shown below, the exact form of this initial density distribution does not play a big role, while the separation of DM particles immediately after their kinetic decoupling is important. After their kinetic decoupling, the DM particles begin to fly apart in the PBH gravitational field, having some velocity distribution (a deformed Maxwell distribution). Some of the particles with low velocities remain gravitationally bound to the PBH, forming subsequently a density spike around it. In the second case, if $M_{\rm BH}>40M_\odot$, such a PBH is formed already after the kinetic decoupling of DM particles, and there is no initial overdensity of radiation and DM around the PBH. In this case, the DM particles with low velocities also remain in finite orbits around the PBH, producing a density spike. Thus, DM density spikes are formed around PBHs at the radiation-dominated stage. After the onset of the matter-dominated stage in the Universe, the DM mass around PBHs begins to grow during the secondary accretion, and a universal density profile $\rho\propto r^{-9/4}$ is formed. The DM density in the central regions of the spikes is so large that by now the DM particles have managed to annihilate (under the assumption that standard neutralinos constitute the DM) at distances that exceed the gravitational PBH radii by several orders of magnitude. For this reason, to calculate the present-day density profile around PBHs, it will be sufficient for us to consider the phenomena at great distances from the PBHs, where Newtonian gravitational dynamics is a good approximation and the general relativity effects are unimportant. The DM remaining at great distances continues to annihilate at present, producing signals in gamma-ray emission. Comparison of the calculated signals with the Fermi-LAT data allows the number of PBHs to be constrained. The annihilation of DM particles in clumps around PBHs has already been considered in \cite{p28,p38,p17}, where constraints on the cosmological PBH density parameter were obtained. Calculations \cite{p28} and \cite{p38} assumed the density profile in the central region of a clump to be close to $\rho\propto r^{-3/2}$, while \cite{p17} considered power-law profiles $\rho\propto r^{-\alpha}$ with $\alpha=1.5-3$. The annihilation of DM in density cusps around black holes was considered in \cite{Sanetal10}, \cite{Sanetal11}, \cite{Sanetal12}, and new gamma-ray constraints were obtained. The goal of this paper is to calculate the density profile in the central region of DM clumps around PBHs by taking into account the initial thermal velocity distribution of DM particles after their kinetic decoupling. We will show that the density profile has a more complex form than $\rho\propto r^{-\alpha}$. Knowledge of the density profile allows one to calculate the signals from DM annihilation around PBHs more reliably and to obtain constraints on the number of PBHs in the Universe.

DM clumps can be produced by various mechanisms \cite{p6}. They can be formed both from cosmological density perturbations in the dark matter itself and around compact seed masses, for example, around cosmic strings \cite{p27} or PBHs \cite{p15,p35,p30,p36,p28,p38}. Secondary accretion, the infall and virialization of cold DM onto PBHs, was thought to be the main mechanism for the formation of DM clumps around PBHs. However, we showed in this paper that there exists another mechanism that gives rise to denser DM clumps around PBHs than was considered in secondary accretion models. The DM density around PBHs grows at the radiation-dominated stage due to the presence of slow DM particles in their thermal velocity distribution. Fairly slow particles after their kinetic decoupling are in finite orbits around PBHs and produce high-density DM clumps. Considering the kinematics of particles around PBHs allowed the density profile to be found. The DM particles in the central regions of clumps have managed to annihilate by now. However, the remaining halos are still very dense, and intense annihilation occurs in them. This effect can be interesting for experiments on indirect detection of DM particles though the search for their annihilation products, because the annihilation of particles in spikes can contribute to the observed gamma-ray emission. Comparison of the calculated signal with the Fermi-LAT observational limits gives upper bounds on the present-day cosmological PBH density parameter from $\Omega_{\rm BH}\leq1$ to $\Omega_{\rm BH}\leq10^{-8}$, depending on the PBH masses at $M_{\rm BH}\geq10^{-8}M_\odot$. Comparable (in magnitude) but weaker constraints $\Omega_{\rm BH}\leq 10^{-4}$ were obtained previously in \cite{p28}, where the density profile in a spike was assumed to be $\rho\propto r^{-3/2}$. However, it should be noted that our constraints are largely model-dependent ones: they depend fundamentally on the as yet unknown properties of DM particles. The derived constraints refer to standard neutralinos or to other DM particles having the properties of weakly interacting massive particles (WIMPs), i.e., having masses and annihilation cross sections comparable to them in order of magnitude. For other DM particles, both the velocity distribution (\ref{maxdistr}) and the annihilation signals can be significantly different. For example, if the DM particles do not annihilate at all, then the formation of a DM density spike around PBHs is still possible, but, in this case, there is no early annihilation and no decrease in central density and the signals are absent in the cosmic gamma-ray emission. The density spikes can be bright gamma-ray sources only at certain masses and annihilation cross sections of DM particles. The derived constraints do not refer, for example, to the models in which the DM consists of PBHs. Nevertheless, the neutralinos in nonminimal supersymmetric models so far remain among the most probable DM candidates, and the constraints obtained here can hold. Author is grateful to V.K.~Dubrovich for useful discussions.

1607.00612

1607

1607.08786_arXiv.txt

In the context of Standard Model Extensions (SMEs), we analyse four general classes of Super Symmetry (SuSy) and Lorentz Symmetry (LoSy) breaking, leading to {observable} imprints at our energy scales. The photon dispersion relations show a non-Maxwellian behaviour for the CPT (Charge-Parity-Time reversal symmetry) odd and even sectors. The group velocities exhibit also a directional dependence with respect to the breaking background vector (odd CPT) or tensor (even CPT). In the former sector, the group velocity may decay following an inverse squared frequency behaviour. Thus, we extract a massive and gauge invariant Carroll-Field-Jackiw photon term in the Lagrangian and show that the mass is proportional to the breaking vector. The latter is estimated by ground measurements and leads to a photon mass upper limit of $10^{-19}$ eV or $2 \times 10^{-55}$ kg and thereby to a potentially measurable delay at low radio frequencies.

1607.08786

1607

1607.08265_arXiv.txt

White dwarfs are compact stars, similar in size to Earth but $\mathbf{\sim 200}$,$\mathbf{000}$ times more massive$^{1}$. Isolated white dwarfs emit most of their power from ultraviolet to near-infrared wavelengths, but when in close orbits with less dense stars, white dwarfs can strip material from their companions, and the resulting mass transfer can generate atomic line$^{2}$ and X-ray$^{3}$ emission, as well as near- and mid-infrared radiation if the white dwarf is magnetic$^{4}$. However, even in binaries, white dwarfs are rarely detected at far-infrared or radio frequencies. Here we report the discovery of a white dwarf / cool star binary that emits from X-ray to radio wavelengths. The star, AR~Scorpii (henceforth AR~Sco), was classified in the early 1970s as a $\delta$-Scuti star$^{5}$, a common variety of periodic variable star. Our observations reveal instead a $\mathbf{3.56}\,$hr period close binary, pulsing in brightness on a period of $\mathbf{1.97}\,$min. The pulses are so intense that AR~Sco's optical flux can increase by a factor of four within $\mathbf{30}\,$s, and they are detectable at radio frequencies, the first such detection for any white dwarf system. They reflect the spin of a magnetic white dwarf which we find to be slowing down on a $\mathbf{10^7}\,$yr timescale. The spin-down power is an order of magnitude larger than that seen in electromagnetic radiation, which, together with an absence of obvious signs of accretion, suggests that AR~Sco is primarily spin-powered. Although the pulsations are driven by the white dwarf's spin, they originate in large part from the cool star. AR~Sco's broad-band spectrum is characteristic of synchrotron radiation, requiring relativistic electrons. These must either originate from near the white dwarf or be generated \textit{in situ} at the M star through direct interaction with the white dwarf's magnetosphere.

1607.08265

1607

1607.03494_arXiv.txt

A homogeneous search for stellar flares has been performed using every available \Kepler light curve. An iterative light curve de-trending approach was used to filter out both astrophysical and systematic variability to detect flares. The flare recovery completeness has also been computed throughout each light curve using artificial flare injection tests, and the tools for this work have been made publicly available. The final sample contains 851,168 candidate flare events recovered above the 68\% completeness threshold, which were detected from 4041 stars, or 1.9\% of the stars in the \Kepler database. The average flare energy detected is $\sim$$10^{35}$ erg. The net fraction of flare stars increases with $g-i$ color, or decreasing stellar mass. For stars in this sample with previously measured rotation periods, the total relative flare luminosity is compared to the Rossby number. A tentative detection of flare activity saturation for low-mass stars with rapid rotation below a Rossby number of $\sim$0.03 is found. A power law decay in flare activity with Rossby number is found with a slope of -1, shallower than typical measurements for X-ray activity decay with Rossby number.

Flares occur on nearly all main sequence stars with outer convective envelopes as a generic result of magnetic reconnection \citep{pettersen1989}. These events occur stochastically, and are most frequently observed on low-mass stars with the deepest convective zones such as M dwarfs. Solar and stellar flares are believed to form via the same mechanism: a magnetic reconnection event that creates a beam of charged particles which impacts the stellar photosphere, generating rapid heating and the emission we observe at nearly all wavelengths. Numerical simulations are now able to describe much of the physics for solar and stellar flares and their effect on a star's atmosphere \citep{allred2015}. Flare occurrence frequency and event energy are connected to the stellar surface magnetic field strength. Reconnection events on the Sun typically occur around a sunspot pair (or bipole) or between a group of spots. Surface magnetic field strength deceases over the life a star, due to a steady loss of angular momentum which quiets the internal dynamo \citep{skumanich1972}. Older, slowly rotating stars like our Sun exhibit smaller and fewer starspots, while young, rapidly rotating stars can produce starspots that are long lived and cover a significant portion of the stellar surface. Flares are known to follow this same basic trend \citep{ambartsumian1975}. For example, young T Tauri systems are known to be highly active with frequent flares \citep{haro1957}. Maximal flare energies have also been proposed as a means for constraining the age of field stars \citep[e.g.][]{parsamyan1976,parsamyan1995}. The duration of a star's life that it produces frequent large spots and flares may dramatically affect planetary, atmospheric, and biological processes, and thus impact planet habitability. This is particularly important for planets around low-mass stars, whose flares can produce extremely high amounts of UV and X-ray flux, and whose active lifetimes are much longer than Solar-type stars \citep{west2008}. To better understand the impact flares might pose for habitability, \citet{segura2010} modeled the affect of a single large stellar flare on a Earth-like planet's atmosphere. For a single large flare this study found only a short timescale increase in biologically harmful UV surface flux, and full planetary atmosphere recovery within two years. However, due to the possibility of repeated flaring and constant quiescent UV emission, concerns remain about UV flux from active and flaring stars, and their impact on planetary atmosphere chemistry \citep{france2014}. Given the variety of possible exoplanetary system configurations, it may also be possible for stellar activity and planetary dynamics to conspire to improve planetary habitability conditions \citep{luger2015}. While the impact flares have to planet habitability is an ongoing topic of research, they pose a clear difficulty in exoplanet detection and characterization \citep{poppenhaeger2015}. Due to their short timescales and stochastic occurrences, generating a complete sample of flares for a single star has been very resource intensive, and has only been accomplished for a handful of active stars. Contrast between flares and the quiescent star is also greatest for cooler stars such as M dwarfs, and has led to fewer flare studies for field G dwarfs. Flare rates for ``inactive'' stars like the Sun are largely unconstrained. However, recent space-based planet hunting missions like \Kepler \citep{borucki2010} have started to collect some of the longest duration and most precise optical light curves to date. These unique datasets are ideal for developing complete surveys of stochastic events like flares from thousands of stars, and have begun to revolutionize the study of stellar flares. For example, \citet{davenport2014b} gathered the largest sample of flares for any single star besides the Sun using 11 months of \Kepler data, and used this homogeneous sample to develop an empirical template for single flare morphology. To help characterize the environments of planets found using \Kepler, \citet{armstrong2016} have investigated the rates of very large flares for 13 stars that host planets near their habitable zones. \citet{maehara2012} have used \Kepler data to show a connection between flare rate and stellar rotation in field G dwarfs, in general agreement with activity--age models. In this paper I present the first automated search for stellar flares from the full \Kepler dataset. The flare event sample generated here is unique in carefully combining both long and short cadence data to accurately measure each star's flare rate over the entire \Kepler mission. I have also performed extensive flare injection tests for multiple portions of each light curve, quantifying the completeness limits for flare recovery over time. I demonstrate the utility of this large sample by comparing the flare activity level with stellar rotation and Rossby number, which reveals a clear connection between flares and the evolution of the stellar dynamo as stars age.

I have presented a homogeneous search for stellar flares using every available light curve from the primary 4-year \Kepler mission. A final sample of 4041 flare stars was recovered, with 851168 flare events having energies above the locally determined completeness limit. This analysis included extensive completeness testing, using artificial flare injection and recovery tests throughout each light curve to determine the flare recovery efficiency as a function of time. While these tests provide a robust and straightforward means to estimate the event recovery efficiency, they currently do not estimate how accurately artificial flare event energies were reproduced. Future improvements to the flare finding algorithm could keep track of the recovered energy and duration for every simulated flare. The light curve de-trending algorithm may also be simplified by using more advanced techniques, such as continuous autoregressive moving average-type models to describe the many forms and timescales of variability at once \citep[e.g.][]{kelly2014}. As a demonstration, in Figure \ref{fig:ffd} I have shown one example of a deviation or break from a single power law in flare occurrence at large flare energies. However, many other active stars show similar breaks at large flare energies in this sample. A systematic follow-up study of FFDs is needed to determine if this break is common among young Solar-type or low-mass stars, which will be impact detailed studies of superflare occurrence. The maximum flare energies recovered in this work are also much higher than previous studies, with a small number of stars in Figure \ref{fig:maxcolor} exhibiting up to $10^{39}$ erg events. These events may be the result of errors in either the light curve de-trending leading to spurious flare events, or the quiescent luminosity determination yielding incorrect energies for real events. Note also that small offsets between flare energies calculated with short- and long-cadence data are seen, as in Figure \ref{fig:ffd}. This may be largely an effect of the respective light curve sampling \citep[e.g. see][]{maehara2015}. From the final sample of 4041 flare stars, 402 were found to have published rotation periods from \citet{mcquillan2014}. A striking evolution of flare activity with stellar Rossby number is seen. This evolution includes a possible saturated flare regime for rapidly rotating (low Rossby number) stars, and power-law decay that is qualitatively similar to previous results for chromospheric H$\alpha$ emission. The tentative discovery of a flare saturation regime gives credence to the model of magnetic activity reaching a peak level due to a maximum filling factor of small scale active regions on the surface \citep{vilhu1984}. However, the Rossby saturation limit (Ro$_{sat}$) and the power-law decay slope do not match expected values from most previous studies of magnetic activity saturation and evolution. Since the sample of flare stars is biased more towards K and M dwarfs than most studies of coronal or chromospheric saturation, the smaller Ro$_{sat}$ value may indicate lower mass stars have different saturation limits than Solar-type stars \citep{west2009}. Alternatively, this result may indicate flare activity traces a fundamentally different component of the stellar surface magnetic field. The connection between white light flares, chromospheric emission, coronal heating, and the generation of the magnetic dynamo clearly deserves further observational investigation. Given the varied dependance on Rossby number that these related manifestations of magnetic activity have shown, the dependence of Rossby number as the fundamental metric for tracing dynamo evolution is uncertain \citep{basri1986,stepien1994}. The large sample of flares observed by \Kepler enables a new generation of statistical studies of magnetic activity. This may yield power advances in constraining stellar ages via flare rates or maximum flare energies, known as ``magnetochronology''. The uniformity of flare activity evolution can be tested using wide binary stars or stellar clusters, many of which are being observed by the \Kepler and K2 missions. Beyond the total flare activity levels for ensembles of stars, the temporal morphology of individual flare events may shed new light on the formation of ``classical'' versus ``complex'', multi-peaked flares, as discussed by \citet{davenport2014b}, \citet{balona2015}, and \citet{davenport2015c}. Modeling the detailed structure of these complex events will help in detecting rare ``quasi-periodic pulsations'' in flares \citep{pugh2015}. Finally, the statistical knowledge we gain from \Kepler will enable more accurate predictions of flare yields from future photometric surveys.

1607.03494

1607

1607.08862_arXiv.txt

Motivated by the discovery of the non-thermal Fermi bubble features both below and above the Galactic plane, we investigate a scenario in which these bubbles are formed through Galacto-centric outflow. Cosmic rays (CR) both diffusing and advecting within a Galactic breeze outflow, interacting with the ambient gas present, give rise to $\gamma$-ray emission, providing an approximately flat surface brightness profile of this emission, as observed. Applying the same outflow profile further out within the disk, the resultant effects on the observable CR spectral properties are determined. A hardening in the spectra due to the competition of advective and diffusive propagation within a particular energy range is noted, even in the limiting case of equal CR diffusion coefficients in the disk and halo. It is postulated that this hardening effect may relate to the observed hardening feature in the CR spectrum at a rigidity of $\approx 200$\,GV.

\label{intro} The presence of a Galactic wind has considerable impact on an array of topics connected to describing the Galactic \lq\lq halo\rq\rq\/ environment. With little knowledge about such outer regions of the Galaxy, information provided by non-thermal probes hold the first clues to revealing new information about this Galactic frontier. Over the past few decades, a growing body of evidence has amounted suggesting that our Galactic center (GC) region feeds a wind. Such indications have been provided from a broad observational energy range, from radio HI~\cite{Lockman:1984}, infrared~\cite{Morris:1996} to X-ray~\cite{Cheng:1996}. Infrared observations at larger scales~\cite{Bland-Hawthorn:2003} have further indicated that this wind continues out to larger scales and may be responsible for the larger out-of-plane scale structures observed. More recently, absorption line features in the spectra of distant AGN have been used to probe the gas flow structure \cite{Keeney:2006au}. The picture provided by these results indicates the presence of coherent gas flow, consistent with that of an outflow directed away from the Galactic plane. Furthermore, recent $\gamma$-ray and radio observations \cite{Su:2010qj,Yang:2014pia,Fermi-LAT:2014sfa,Carretti:2013sc} of the region above and below the GC indicate the presence of extended non-thermal particle populations inside bubble structures which sit above and below the Galactic disk. The presence of these cosmic ray (CR) populations are indicative of outflow activity from the GC region. The present picture, therefore, appears to indicate that both hot gas and non-thermal particles are conveyed out from the center of the disk into the halo within a centrally driven Galactic wind. With regards the velocity of the Milky Way's outflow, there are several indicators about this from a host of independent observations. Relatively mild velocities $\sim 300$\,km\,s$^{-1}$ are suggested to be present in the outflow region close to the disk ($\sim 1-2$\,kpc) by the weakness of the X-ray features associated with the bubble edges \cite{Su:2010qj,Kataoka:2013tma,Fang:2014hea,Fox:2015}. The observation of high velocity clouds in regions consistent with the bubble's location \cite{Keeney:2006au}, motivate outflow velocities of $\sim 150$\,km\,s$^{-1}$ at distances of $\sim 4$\,kpc and $\sim 9$\,kpc away of the Galactic plane. Further out towards the edges of the bubbles, other indications support velocities $<100$\,km\,s$^{-1}$ in the outflow. Within such a profile scheme, the distortion of the outflow structures seen to high latitudes in radio observations \cite{Carretti:2013sc} may be related to the motion of the Milky Way towards Andromeda, whose relative velocity is $\sim 50$\,km\,s$^{-1}$. In the following, we consider the secondary signatures that CR embedded in outflows can produce. In Section~\ref{GC_outflow} we adopt simple descriptions for the velocity flow in the outflow and consider the subsequent diffusive-advective motion of CR within it. The generation of secondary signals by these CRs are considered in an effort for simple comparisons with recent observations. In Section~\ref{CR_at_Earth}, the implications of the presence of Galactic driven outflows on the CR detected at Earth are considered. We draw our conclusions from these results in Section~\ref{conclusion}.

\label{conclusion} We first investigated a scenario in which an advective outflow, emanating from the Galactic center region, carries pre-accelerated CR. These CR produce secondary $\gamma$-rays via $pp$ interactions on target gas. We have demonstrated that one can reproduce a flat $\gamma$-ray surface brightness profile, as is observed for the Fermi bubbles, provided that the outflow decelerates with distance above the Galactic disk. Such a description for the outflow profile is encapsulated by \lq\lq breeze\rq\rq\/ solutions of isothermal winds. Assuming CR propagation beyond the central zone is purely diffusive, it is possible for a non-negligible fraction of CR from the GC region to reach large radii. The contamination under this assumption is energy dependent, and we found that CR from the GC may potentially become the dominant source for the flux observed at Earth, at $\gtrsim$~PeV energies. The absence of evidence indicating the onset of a new component in the CR spectrum at these energies, however, place challenges on such a possibility. Imposing, instead, a wind scenario also out at larger Galactocentric radii, we have demonstrated that for the breeze profile~(\ref{EqnWind}), an inflection point is introduced into the CR spectrum shape at $z=0$, as a result of competition between CR advection and diffusion in the halo. We have shown that hardenings can appear in the CR spectrum due to the launching of a breeze or wind in the Milky Way's halo, even without invoking any change in the CR diffusion coefficient value between the disk and halo. We conclude that a breeze outflow scenario from the Galaxy provides an interesting array of observational signatures able to diagnose its presence. Although presently only motivated from Galactocentric outflow observations, the results outlined provide a useful reference for future observations able to disclose its presence at larger radii.

1607.08862

1607

1607.01943_arXiv.txt

{ Outflows of photoionized gas are commonly detected in the X-ray spectra of Seyfert 1 galaxies. However, the evidence for this phenomenon in broad line radio galaxies, which are analogous to Seyfert 1 galaxies in the radio-loud regime, has so far been scarce. Here, we present the analysis of the X-ray absorption in the radio-loud quasar 4C +74.26. With the aim of characterizing the kinetic and the ionization conditions of the absorbing material, we fitted jointly the XMM-\textit{Newton} Reflection Grating Spectrometer (RGS) and the \textit{Chandra} High Energy Transmission Grating Spectrometer (HETGS) spectra, which were taken 4 months apart. The intrinsic continuum flux did not vary significantly during this time lapse. The spectrum shows the absorption signatures (e.g., Fe-UTA, \ion{O}{vii}, and \ion{Ne}{vii}--\ion{Ne}{x}) of a photoionized gas outflow ($\nh \sim 3.5 \times 10^{21}$ \colc, $\log \xi \sim 2.6$, $v_{\rm out}\sim 3600$ \kms) located at the redshift of source. We estimate that the gas is located outside the broad line region but within the boundaries of the putative torus. This ionized absorber is consistent with the X-ray counterpart of a polar scattering outflow reported in the optical band for this source. The kinetic luminosity carried by the outflow is insufficient to produce a significant feedback is this quasar. Finally, we show that the heavy soft X-ray absorption that was noticed in the past for this source arises mostly in the Galactic ISM.}

\label{intro5} In the last fifteen years, thanks to the advent of high-resolution X-ray spectrometers, such as the \xmm \th \th Reflection Grating Spectrometer (RGS) or the \chandra Low and High Energy Transmission Grating Spectrometers (LETGS and HETGS), our knowledge of the circumnuclear gaseous environment of active galactic nuclei (AGN) has advanced significantly.\\ It is now established that roughly half of all local Seyfert galaxies host a photoionized warm absorber (WA) that produces features detectable in the X-ray and in the UV band \citep{cre2003}. These absorption lines are usually blueshifted with respect of the systemic velocity, which indicates a global outflow of the absorbing gas. Spectroscopical observations allow the physical conditions (kinematics and ionization) of the gas to be characterized with high accuracy \citep[see][for a review]{cos2010}. In photoionization equilibrium, the ionization parameter $\xi=L_{\rm ion}/ n r^2$ (where $L_{\rm ion}$ is the ionizing luminosity between 1 and 1000 Ryd, $n$ is the gas density, and $r$ is the distance from the ionizing source) parameterizes the state of the gas. In the X-ray band, there are many transitions, for example from ionized C, N, O, Ne, and Fe, which allow an accurate solution for $\xi$ to be determined. From spectroscopical observables, useful constraints can be put on the gas location \citep{blu2005}; these constraints quantify how much momentum is transferred by the outflow to the surrounding medium \citep[e.g.,][]{cre2012}.\\ The studies of WA in Seyfert galaxies show that these outflows span roughly four orders of magnitude in ionization ($ \log \xi \sim 0 - 4$) and reach velocities of a few thousand \kms \citep{mck2007}. They are often located as far as the putative torus \citep{blu2005}. Some outliers may be located closer to the nucleus, at the distance of the accretion disk or farther out in the galaxy at $\sim$kpc distance from the center \citep[][]{dig2013}. In most of the cases, the kinetic power of the WA is found to be negligible with respect to the AGN radiative power \citep[e.g.,][]{ebr2016}. Thus, WA are not expected to play a significant role in a possible negative AGN feedback \citep{sca2004,som2008, hop2008,hop2010,kin2015}. \\ A different class of photoionized winds are the ultrafast outflows (UFO). These may be present in 35\% of Seyfert galaxies \citep{tom2010} and differ from classical WA because of the higher outflow velocity (v$\sim$0.1 c, where c is the speed of light) and of the higher ionization ($\log \xi \geq 3 $, \citealt{tom2011}). Hence, because of the higher energy and higher blueshift of their transitions (e.g., \ion{Fe}{xxv}--\ion{Fe}{xxvi}), UFO are detectable only in lower resolution CCD spectra. These powerful winds are believed to be a nuclear phenomenon originating from the accretion disk \citep{tom2012, nar2015}.\\ The detection of photoionized features in broad line radio galaxies (BLRG), which are analogous to Seyfert 1 galaxies in the radio-loud regime, was expected to be difficult because of the presence of a relativistic jet. The Doppler-boosted, non-thermal radiation of a jet located close to the line of sight could actually mask the absorption features. So far, the statistics of known WA in BLRG relies on a handful of cases, of which only three are WA detections in a high-resolution X-ray dataset.\\ Hints of photoionized absorption were noticed, for instance, in the ROSAT-PSPC spectrum of 3C 351 \citep{fio1993} and 3C 212 \citep{3c212}. Interestingly, these two sources also display WA features in the UV \citep{math1994, yua2002}. More recently, \citet{mol2015} have reported the detection of \ion{O}{vii} and of \ion{Fe}{xx} absorption edges in the EPIC-pn spectrum of IGR J14488-4008, a giant radio-loud galaxy discovered by INTEGRAL.\\ The first case of a WA in a BLRG studied with a grating spectrum was a long \heg spectrum of 3C 382 \citep{ree2009}. The detection of this WA, whose location is consistent with the distance of the narrow line region (NLR), was promptly confirmed by a subsequent RGS observation \citep{tor2010}. A second case is the remarkable photoionized outflow in 3C 445. In the \heg spectrum of this source \citet{ree2010} detected a low-ionization outflow moving at a sub-relativistic velocity. A deep Suzaku spectrum also shows indications of blueshifted absorption from highly ionized iron \citep{bra2011}. Both these spectra are consistent with a scenario where our line of sight intercepts an equatorial disk wind located at $\sim$sub-pc scale. The low-ionization absorber may consist of sparse clumps embedded in a highly ionized wind \citep{ree2010}. In addition to these two cases, \citet{tor2012} report a WA detection in the RGS spectrum of 3C 390.3.\\ Signatures of more highly ionized UFO have also been detected in the CCD spectra of a handful of radio-loud sources \citep{tom2014}, with a statistical incidence comparable, within the uncertainties, to what is found for radio-quiet Seyferts.\\ \\ In this paper we present the analysis of the X-ray grating spectrum -- obtained with the RGS and the \chandra-HETGS -- of the BLRG 4C +74.26. This source is located at a redshift of 0.104 \citep{ril1988}. In the optical, it shows broad permitted lines with a a \fwhm \th of 10\,000 \kms\ for the H$_{\rm \beta}$ line \citep{win2010}. Using this line width a SMBH mass of $3 \times 10^{9}$ \msun\ \th is inferred.\\ Because of its $\sim$1 Mpc projected linear size \citep{ril1988}, this source is the largest known radio source associated with a quasar. Its radio morphology is typical for a Fanaroff-Riley type II source (FRII), although the 178 MHz radio luminosity is borderline with the type I class (FRI). Observations with the Very Large Array (VLA) have revealed a one-sided jet which is at least 4 kpc long \citep{ril1990}. The flux limit for a counter-jet, which could be set with a subsequent Very Long Baseline Interferometry (VLBI) observation \citep{pea1992}, implies that the source axis lies at $\la 49^{\circ}$ from our line of sight.\\ Evidence of a high-velocity outflow in \qc was found in the optical spectropolarimetric analysis performed in \citet{rob1999}. These authors noticed that the broad H$\alpha$ line appears redshifted in polarized light, which can be explained if the scattering medium producing the polarization is part of a polar outflow. \\ Since 1993, \qc has been targeted by many X-ray observatories, including ROSAT, ASCA, \textit{Beppo}-SAX \xmm, and \textit{Suzaku}. In the \xmm \th \th \citep{bal2005} and in the Suzaku \citep{lar2008} spectrum a broadened \fek \th emission line has been clearly detected at 6.4 keV. Recently, in the Suzaku \citep{gof2013} and the \xmm \th \th spectrum \citep{tom2014} additional absorption features in the Fe-K band have been noticed. These could be due to a highly ionized UFO, with a measured outflow velocity on the order of $\sim0.1c$.\\ By studying the correlations between the Suzaku light curves in different bands, \citet{nod2013} were able to extract the stable soft-excess component \citep{sin1985} that may dominate the continuum emission at soft energies (i.e., below 2.0 keV). According to these authors, the most likely origin for the soft-excess in this source is thermal Comptonization of the disk photons in a warm plasma \citep[see, e.g.,][]{nod2011, don2012, jin2012, pet2013, dig2014, giu2015, boi2016}.\\ It was found, however, that the soft-excess underlies a heavy soft X-ray absorption. For instance, absorption from a substantial column density of gas in excess at the Galactic column density was seen earlier on in the ROSAT-PSPC \citep{bri1998}, ASCA \citep{bri1998,sam1999,ree2000}, and Beppo-SAX \citep{has2002} spectra. In a more recent \xmm \th \th observation, \citet{bal2005a} detected a column of cold absorption greater than the Galactic value, with an intrinsic column of $\sim 1.9 \times 10^{21}$ \colc. Moreover, the broadband \xmm \th \th spectrum shows evidence of a weak WA intrinsic to the source. The WA is highlighted by features identified as the \ion{O}{vii} and \ion{O}{viii} absorption edges \citep{bal2005}.\\ Motivated by these indications of a complex absorption in this source, here we use the archival \xmm \th \th Reflection Grating Spectrometer (RGS) and \chandra High Energy Trasmission Grating Spectrometer (HETGS) spectra of 4C +74.26 to characterize for the first time the kinematics and the ionization condition of the X-ray absorbing material.\\ In Sect. \ref{dataqc} we describe our data reduction procedure. Then in Sect. \ref{sedqc} we build the spectral energy distribution (SED), and in Sect. \ref{specqc} we perform the spectral analysis. Finally, in Sect. \ref{dsou} we discuss our results, and in Sect. \ref{concqc} we state the conclusions.\\ The C-statistic \citep{cas1979} is used throughout the paper, and errors are quoted at the 68$\%$ confidence level ($\Delta C=1.0$). In all the spectral models presented, we use the total Galactic hydrogen column density from \citet[][$\nh=2.31 \times 10^{21}$ \colc]{wilh22013}. In our luminosity calculations we use a cosmological redshift of z=0.104 and a flat cosmology with the following parameters: \ho=70 \kmsmpc, \omegam=0.3, and \omegalambda=0.7. \begin{table*} \caption{XMM-Newton and \chandra observation log for 4C +74.26.} \label{obs5.tab} \centering \begin{tabular}{lccccc} \hline\hline Date & Instrument & Observation ID & Net exposure \tablefootmark{a} & F$_{\rm 0.3-2.0 \, keV}$\tablefootmark{b} & F$_{\rm 2.0-10.0 \, keV}$\tablefootmark{b} \\ & & & (ks) & \multicolumn{2}{c}{($10^{-11}$ \ergsc)} \\ \hline 2003 Oct 6 & HETGS & 4000 & 37 & 0.7 & 2.8 \\ 2003 Oct 8 & HETGS & 5195 & 31 & 0.8 & 2.9 \\ 2004 Feb 6 & RGS & 0200910201 & 34 & 0.9 & 3.0 \\ \hline \end{tabular} \tablefoot{ \tablefoottext{a}{Resulting exposure time after correction for background flares.} \tablefoottext{b}{Observed flux in the quoted bands.} } \end{table*} \begin{figure}[t] \includegraphics[angle=180,width=0.5\textwidth]{sedqc.pdf} \caption{Spectral energy distributions for 4C +74.26. Filled circles: OM fluxes corrected for the Galactic extinction. Open squares: X-ray intrinsic continuum obtained from a phenomenological fit of the EPIC-pn spectrum.} \label{sedqc.fig} \end{figure}

\label{dsou} We have presented a joint analysis of the RGS and HETGS spectra of the heavily X-ray absorbed radio-loud quasar 4C +74.26. Thanks to the high spectral resolution of these grating spectra, we were able to reveal a rich spectrum of absorption features originating from both Galactic and intrinsic material. In our analysis we used the total Galactic column density given in \citet{wilh22013}, which includes the contribution of molecular hydrogen. This is roughly twice the value provided by 21 cm surveys. The enhanced Galactic absorption explains the heavy suppression of the soft X-ray flux that was noticed in the past for this source \citep{bri1998, sam1999,ree2000,has2002, bal2005}.\\ The intrinsic absorption comprises a highly ionized WA which produces a deep Fe-UTA trough in the RGS and the weak absorption features that are visible in the HETGS spectrum. We found that an outflow velocity of $\sim 3600$ \kms\ is required for a best fit of the absorption features visible in the two spectra. This finding is evidence for WA absorption in radio-loud objects, which so far has been scarce. Indeed, in addition to 3C 382, 3C 445 and 3C 390.3, \qc is the fourth radio-loud source where a photoionized outflow has been clearly characterized in a high-resolution dataset. The column density, ionization parameter, and outflow velocity that we measured for the WA in 4C +74.26 are within the range observed in Seyfert 1 galaxies \citep{mck2007} and are also in line with the values found in 3C 382 and 3C 390.3, the other two radio-loud galaxies hosting a classical WA. The case of 3C 445 is an outlier, as this source hosts a high-velocity, high-column UFO-like wind \citep[see the review of][]{tor2012}.\\ In the following sections we use the results of our spectral analysis and the information from the literature to infer a possible geometrical model for the outflow is this AGN. To this purpose, in Sect. \ref{dion} we estimate the possible location and the energetics of the warm absorber. In Table \ref{disc.tab}, upper panel, we outline some basic physical properties of the source that serve for an order of magnitude comparison. We took the black hole mass $M_{\rm BH}$ and the source inclination $i$ from the literature as already explained in Sect. \ref{intro5}. From a numerical integration of the SED of Fig. \ref{sedqc.fig} we computed the ionizing luminosity $L_{\rm ION}$ between 1 and 1000 Ry and the bolometric luminosity over the whole optical and X-ray band. We note that the bolometric luminosity is probably underestimated because the radio emission at low energies and the gamma ray emission at high energies are not included in our SED. Hence, using these data we estimated the Eddington luminosity $L_{\rm Edd}$ and the mass accretion rate $\dot{M}_{\rm acc}$, for which we assumed an accretion efficiency $\eta=0.1$. For the jet power $P_{\rm jet}$ we used the radio flux at 1.4 GHz \citep{con1998} and the scaling relationship of \citet{cav2010}. The radius of the broad line region $R_{\rm BLR}$ scales with the optical luminosity at 5100 \AA\ \citep{wan2002}. The luminosity value is given in \citet{win2010}. Finally, the radius of the putative torus $R_{\rm TOR}$, which is nominally set by the dust sublimation radius, scales with $L_{\rm ion}$ \citep{kro2001}. \subsection{Location and energetics of the ionized outflow} \label{dion} In Table \ref{disc.tab}, lower panel, we outline some physical properties of the ionized outflow that we estimated using our measured parameters, namely $\nh\sim 3.1\times 10^{21}$ \colc, $\log \xi \sim 2.6$, and $v_{\rm out}\sim 3600$ \kms. We follow here the argumentation of \citet{blu2005}, which assumes that the outflow is a partially filled spherical shell of gas with a volume filling factor $f$. An analytical expression for the volume filling factor $f$ is derived in \citet{blu2005} from the prescription that the kinetic momentum of the outflow must be on the order of the momentum of the absorbed radiation plus the momentum of the scattered radiation. For the ionized outflow in \qc we found that the ionized gas fills only $\sim 0.007 \%$ of the spherical volume, which suggests that it may consist of sparse clumps.\\ We set a range of possible distances for the absorber from the conditions that the velocity of the outflow must exceed the escape velocity from the AGN and that the outflowing shell must not be thicker than its distance from the center ($\Delta r / R \leq 1$). Analytically, $$ \frac{2GM_{\rm BH}}{v^2} \leq R \leq \frac{L_{\rm ion} f}{\xi \nh},$$ where G is the gravitational constant. For our parameters, both these expressions return a value of $\sim$2 pc (Table \ref{disc.tab}). This constrains the ionized outflow of \qc to be located outside the BLR ($R_{\rm BLR}=0.2$ pc), but within the boundary of the putative torus ($R_{\rm TOR}=6$ pc). \\ A patchy ionized outflow located outside the BLR is a natural candidate for being the scattering outflow that is required in the \citet{rob1999} analysis of the polarized optical spectrum of this source. Their model prescribes that the observed redshift of the polarized H$\alpha$ line is due to a high-velocity motion of the scattering material which polarizes the BLR light. In this framework, the outflow velocity inferred for the scatterer depends on the inclination of the scattering cone with respect to the jet axis. For the case of a scattering outflow coaligned with the radio jet, they quote a velocity of $\sim$5000 \kms. Interestingly, if we consider the same source inclination used in the \citet{rob1999} model ($\sim 45^{\circ}$) and we assume that the WA found in our analysis is outflowing along the polar axis of the source, we obtain a deprojected velocity of $v_{\rm out}/\cos 45^ {\circ}\sim$5000 \kms\ (Fig. \ref{geo_qc.fig}). This matches the \citet{rob1999} prediction. This correspondence hints at the possibility that the WA detected here and the outflowing polar scatterer discovered in \citet{rob1999} are one and the same.\\ Given the velocity, the mass outflow rate is given by $$ \dot{M}_{\rm out}= \frac{1.23 m_{\rm p} L_{\rm ion} f v_{out} \Omega}{\xi}, $$ where $m_p$ is the proton mass and $\Omega$ is the solid angle of the outflow, which we set to 2.1 sr, as in \citet{tor2012}. This is derived assuming that at least 50\% of radio-loud objects host an outflow, like in a Seyferts galaxy, and using the information that $\sim$33\% of the radio galaxies belonging to the 3CR sample are type 1 AGN \citep{but2009}. Hence, using the mass outflow rate, the kinetic luminosity of the outflow is readily computed as $L_{\rm kin}=\frac{1}{2} \dot{M}_{\rm out} v_{\rm out}^2 $.\\ The value we obtained for the kinetic luminosity is at least four orders of magnitude lower than the bolometric luminosity. Theoretical AGN feedback models \citep[e.g.,][]{dim2005, hop2010} typically require kinetic luminosities comparable with the bolometric luminosity for an outflow to be able to halt the star formation in a typical galactic bulge. Thus, this outflow is unable to deliver a significant feedback in this AGN. Moreover, as found for the other radio-loud galaxies hosting a WA, the kinetic luminosity of the outflow is negligible compared to the jet power ( $L_{\rm kin} \sim 10^{-2} P_{\rm jet}$). Thus, the case of 4C +74.26 confirms that the jet is a more likely driver of AGN feedback in radio-loud galaxies \citep{tor2012}. \\ \begin{table} \caption{Properties of 4C +74.26.} \label{disc.tab} \centering \begin{tabular}{lc} \hline\hline Source properties & Ref \\ \hline $M_{\rm BH}= 3 \times 10^{9}$ \msun & \citet{win2010} \\ $i \leq 49^{\circ}$ & \citet{pea1992}\\ $L_{\rm bol}=9.7 \times 10^{46}$ \ergs & Sect. \ref{dsou} \\ $L_{\rm bol}/L_{\rm Edd}=0.25$ &Sect. \ref{dsou} \\ $\dot{M}_{\rm acc}=17$ \msunyr & Sect. \ref{dsou}\\ $L_{\rm ion}=8.8 \times 10^{46}$ \ergs & Sect. \ref{dsou}\\ $P_{\rm jet}=2 \times 10^{44}$ \ergs &Sect. \ref{dsou} \\ $P_{\rm jet}/L_{\rm Edd}=6 \times 10^{-4}$ & Sect. \ref{dsou}\\ $R_{\rm BLR}=0.2$ pc & Sect. \ref{dsou}\\ $R_{\rm TOR}=6$ pc & Sect. \ref{dsou}\\ \hline Ionized outflow properties & Ref\\ \hline 1.6 $ \leq R \leq 1.8$ pc& Sect. \ref{dion}\\ $f=7 \times 10^{-5}$ & Sect. \ref{dion}\\ $\dot{M}_{\rm out}=0.4$ \msunyr & Sect. \ref{dion}\\ $L_{\rm kin}=1.5 \times 10^{42}$ & Sect. \ref{dion}\\ \hline \end{tabular} \end{table} \begin{figure}[t] \includegraphics[width=0.5\textwidth]{geo_qc.jpg} \caption{Outflow in the inner region of 4C +74.26. The observer's line of sight lies at 45$^{\circ}$ from the jet axis. The WA is part of a polar outflow located outside the BLR. The ionized gas outflows along the polar direction with a velocity of $\sim$5000 \kms, which is observed as $\sim$3500 \kms\ from the observer's inclination angle.} \label{geo_qc.fig} \end{figure}

1607.01943

1607

1607.07879_arXiv.txt

We explore the detection limits of the phase modulation (PM) method of finding binary systems among multi-periodic pulsating stars. The method is an attractive way of finding non-transiting planets in the habitable zones of intermediate mass stars, whose rapid rotation inhibits detections via the radial velocity (RV) method. While oscillation amplitudes of a few mmag are required to find planets, many $\delta$\,Scuti stars have these amplitudes. In sub-optimal cases where the signal-to-noise of the oscillations is lower, low-mass brown dwarfs ($\sim$13\,M$_{\rm Jup}$) are detectable at orbital periods longer than about 1\,yr, and the lowest mass main-sequence stars (0.1--0.2\,M$_{\odot}$) are detectable at all orbital periods where the PM method can be applied. We use purpose-written Markov chain Monte Carlo (MCMC) software for the calculation of the PM orbits, which offers robust uncertainties for comparison with RV solutions. Using \textit{Kepler} data and ground-based RVs, we verify that these two methods are in agreement, even at short orbital periods where the PM method undersamples the orbit. We develop new theory to account for the undersampling of the time delays, which is also necessary for the inclusion of RVs as observational data in the MCMC software. We show that combining RVs with time delays substantially refines the orbits because of the complementarity of working in both the spatial (PM) and velocity (RV) domains simultaneously. Software outputs were tested through an extensive hare and hounds exercise, covering a wide range of orbital configurations including binaries containing two pulsators.

Two methods are traditionally used to detect and characterise binary companions to stars: radial velocities (RVs) from spectroscopy and eclipse measurements from photometry. The availability of four years of high-precision photometry from the \textit{Kepler} Mission has facilitated a third method, namely measuring the effect of binary motion on stellar pulsations. This can be done via the frequency modulation (FM) method \citep{shibahashi&kurtz2012,shibahashietal2015,kurtzetal2015a}, or the phase modulation (PM) method \citep{murphyetal2014,koen2014,balona2014b}. Here we discuss the latter, which uses periodic phase shifts in stellar pulsations to infer a binary companion. Orbits are characterised using the time delays (sometimes called R\o{}mer delays) in place of RVs \citep{teltingetal2012}. Time delays have already been used to detect planetary companions to pulsating subdwarf B stars \citep{silvottietal2007}, building upon the traditional $O-C$ techniques that have been particularly successful with pulsars \citep{wolszczan&frail1992}. We have found the PM method to work well for $\delta$\,Scuti pulsating stars using \textit{Kepler} photometry. All the aforementioned methods have advantages and disadvantages and are therefore complementary. RVs can reveal companions down to planetary masses in stars that have sharp spectral lines, but require a lot of observing time. Eclipses and transits push down to small planetary companions, provided the inclination is favourable, but geometry limits the number of systems observed to eclipse to about 1\:per\:cent. The PM method works best at longer orbital periods where the time delays are larger, but is restricted to stars with stable pulsations. This is the fourth in a series of papers dedicated to development of the PM method. The first \citep{murphyetal2014} described the principle of obtaining the time delays from observed phase shifts of the stellar pulsations. The second \citep{murphy&shibahashi2015} provided an analytical method for fully solving the orbit, even in highly eccentric cases. The methodology contained within those papers is summarised in Sect.\,\ref{ssec:hound}, along with a description of new Markov chain Monte Carlo (MCMC) software used to determine the orbital parameters. A third study was recently made by \citet{comptonetal2016} to determine which kind of oscillating stars are suitable for PM analyses. They found that $\delta$\,Sct stars and white dwarfs were most favourable (cf. \citealt{dalessioetal2015}). However, we note that only around 20 white dwarfs were observed by Kepler during the main mission, and 14 of those were non-pulsators \citep{maozetal2015}, while thousands of $\delta$\,Sct stars were observed \citep{murphy2014}. In this paper we use a hare-and-hounds exercise to investigate the sensitivity of the PM method to various orbits. The individual roles of the hare and the hound are described in Sect.\,\ref{ssec:hare} and \ref{ssec:hound}, respectively. Particular attention has been paid to recovering the orbital parameters for undersampled orbits (Sect.\,\ref{ssec:undersampling}). In addition, we conducted specific experiments to determine the detection limits of the PM method, in terms of both companion mass and maximum orbital period that can be analysed, and the influence of the pulsation properties on these limits (Sect.\,\ref{sec:experiments}). The use of RVs as observational inputs alongside time delays is discussed in Sect.\,\ref{sec:rv} and applied to real \textit{Kepler} data, including a binary system in which both stars pulsate.

We have developed MCMC software to solve binary orbits based on a series of time delay observations, and to provide robust uncertainties. We simulated orbits covering a range of parameters to explore the sensitivity limits of the method, the factors governing those limits, and to predict the lowest mass companions detectable by the method. We confirmed that the method is much more sensitive to stars oscillating with high signal-to-noise, and in such cases the detection limit approaches 1--2\,M$_{\rm Jup}$ at long orbital periods ($>1000$\,d), where the habitable zones of intermediate-mass stars are located. We also showed that orbital solutions can be obtained when the orbital period is longer than the data timespan, with the upper limit on orbital period depending on the orientation and eccentricity of the orbit, and whether or not the periastron phase is observed. The uncertainties on such orbits tend to be large and perhaps underestimated. However, since any overestimates of the orbital period will be correlated with overestimates of $a_1 \sin i / c$, and vice-versa, the mass function is well recovered. This is because the rate of change of the time delays, which is governed by the mass function, can be established without observing a full orbit. One drawback to the PM method is that we must divide the light curve into segments of several days to make adequate measurements of the pulsation phases. For short-period binaries this leads to significant undersampling and if the orbits are eccentric, the time-delay curve is heavily smeared. We have overcome this drawback by developing correction factors to the time-delay fitting function, and we verified the validity and implementation of that function in the MCMC algorithm via a hare and hounds exercise. The undersampling correction can help in understanding our completeness in surveys for binary stars with short periods, but a full completeness analysis remains as future work. Another development is the simultaneous use of radial velocities and time delays as input data for solving the orbits in the MCMC framework. This requires implementation of an undersampling correction the time-delay fitting function, but allows the orbital parameters to be determined much more precisely. This is partly because of the complementarity of the PM and RV methods: the latter is the time derivative of the former and measurement of both provides a clear improvement in constraints on the orbit. Additionally, any RV measurements made now will double the time span of the observations, constraining the orbital period much more tightly, and thereby reducing the uncertainties on the other orbital parameters. For a real \textit{Kepler} binary system we showed that the combination of RVs and time delays can constrain the eccentricity to a factor 50 better than either the time delays or RVs alone, and orbital periods can be measured with a precision of seconds, even without eclipses.

1607.07879

1607

1607.00032_arXiv.txt

We perform a comparative analysis of constraints on sterile neutrinos from the Planck experiment and from current and future neutrino oscillation experiments (MINOS, IceCube, SBN). For the first time, we express joint constraints on $\Neff$ and $\meff$ from the CMB in the $\Delta m^{2}$, $\sin^{2} 2 \theta$ parameter space used by oscillation experiments. We also show constraints from oscillation experiments in the $\Neff$, $\meff$ cosmology parameter space. In a model with a single sterile neutrino species and using standard assumptions, we find that the Planck 2015 data and the oscillation experiments measuring muon-neutrino ($\nu_{\mu}$) disappearance have similar sensitivity.

\label{intro} The search for low-mass sterile neutrinos is motivated by several experimental anomalies that are not consistent with the three-flavour paradigm. Sterile neutrinos would change the oscillation probabilities observed by detecting neutrinos from accelerators, nuclear reactors, or produced in the atmosphere. On a cosmological scale, they would modify the power spectrum of the Cosmic Microwave Background (CMB) (Fig. \ref{fig:cmb_oscillations}). Both types of measurement put severe constraints on the existence of extra neutrino flavours, but they are evaluated in terms of different parameter sets. The CMB measurements constrain the effective number of additional neutrino species, $\DNeff$ (above the Standard Model (SM) prediction of $N_{\rm eff} = 3.046$), and the effective sterile neutrino mass $\meff$. Oscillation experiments parameterize their constraints in terms of mass-squared differences, $\Delta m^{2}_{ij}$, between the mass eigenstates, and the mixing angles $\theta_{\alpha\beta}$ between mass and flavour eigenstates. Here, we use the calculation of~\cite{Hannestad2012} and show the Planck CMB cosmology constraints in the same parameter space as used for $\nu_{\mu}$ disappearance measurements. Several experimental anomalies related to the appearance and disappearance of $\nu_e$ could be explained by light sterile neutrinos with a mass-squared difference relative to the active states of $\Delta m^{2} \approx \unit[1]{eV^{2}}$~\cite{Abazajian2012,Collin:2016rao,Collin2016_globalfit}. The LSND Collaboration observes an excess of $\bar{\nu}_{e}$ appearance in a $\bar{\nu}_{\mu}$ beam~\cite{LSND}, and MiniBooNE measures an excess of both $\nu_e$~\cite{Miniboone2007} and $\bar{\nu}_{e}$ appearance~\cite{Miniboone2010,Miniboone2013}. Reactor experiments observe a deficit of $\approx 6\%$ in the $\bar{\nu}_{e}$ flux compared to expectations~\cite{reactor}. Furthermore, Gallium experiments observe a smaller $\nu_e+\mbox{$^{71}$Ga}\to \mbox{$^{71}$Ge}+e^{-}$ event rate than expected from $^{51}$Cr and $^{37}$Ar sources~\cite{Giunti:2010zu}. The Daya Bay Reactor experiment has searched for $\bar{\nu}_e$ disappearance setting limits on the mixing angle $\sin^2\theta_{14}$ in the low $\Delta m^{2}$ region $0.0002<\Delta m_{41}^{2}< \unit[0.2]{eV^{2}}$~\cite{An:2016luf}. These results have been combined with $\nu_{\mu}$ disappearances searches by MINOS~\cite{MINOS:2016viw} to obtain stringent constraints on the product $\sin^22\theta_{14}\sin^2\theta_{24}$~\cite{Adamson:2016jku}. For this analysis, we focus on recent $\nu_{\mu}$ disappearance results, where no anomalies have been found, and assume that $\sin^2\theta_{14}= \sin^2\theta_{34}=0$ in order to be consistent with the assumptions that were used for deriving these limits. Several studies have combined oscillation and cosmological data to constrain sterile neutrinos. Several~\cite{ref:Archidiacono2012,ref:Archidiacono2013,Gariazzo2013,Archidiacono2014,Archidiacono2016} use the posterior probability distribution on $\Delta m^{2}$ from short-baseline anomalies as a prior in the cosmological analysis. Here, we convert the full CMB cosmology constraints into the oscillation parameterisation and vise versa, focusing on recent $\nu_{\mu}$ disappearance results. This conversion has also been studied in~\cite{Mirizzi2013,Melchiorri2009}. Our analysis differs in several ways: (i) unlike~\cite{Mirizzi2013} we use the 2D combined constraints on $\DNeff$ and $\meff$ in the cosmological analysis, rather than converting 1D constraint values in each parameter individually; (ii) we use the latest CMB data from Planck, updating from the WMAP 5-year data used in~\cite{Melchiorri2009}; (iii) we solve the full quantum kinetic equations, rather than using the averaged momentum approximation~\cite{Mirizzi2012} used in~\cite{Mirizzi2013,Melchiorri2009}; (iv) we also consider the impact of non-zero lepton asymmetry, $L$, and a different sterile mass mechanism. The lepton asymmetry is defined as ${L=(n_{f}-n_{\bar f})N_{f}/N_{\gamma}}$, where $n_{f}$ and $n_{\bar f}$ are the number densities of fermions and anti-fermions, respectively, and $N_{f}$ and $N_{\gamma}$ are the numbers of fermions and photons.

\label{sec:conclusions} In conclusion, we compare sterile neutrino constraints from oscillation experiments and cosmological constraints. We use the quantum kinetic equations to convert between the standard oscillation parameterization of neutrinos (the mass-squared difference and mixing angle) and the cosmology parameterization (the effective sterile neutrino mass and the effective number of neutrino species). We show the relationship between each of the parameter combinations. We show the Planck 2015 CMB cosmology constraints in the oscillation parameter space and find that they rule out large values of $\Delta m_{41}^{2}$ and mixing angle $\theta$. For the fiducial case, the region of parameter space ruled out by IceCube data is already excluded by the Planck CMB constraints. For the first time, we show that much of the MINOS exclusion region is also ruled out by Planck CMB constraints, although for low $\Delta m_{41}^{2}$ MINOS is more constraining. The forecast constraints for the SBN experiments are not expected to add to the information already provided by Planck CMB results with these model assumptions. However, their main sensitivity will be through the $\nu_e$ appearance searches not considered here. The MINOS data adds the most information to that provided by Planck CMB measurements because it probes the lowest $\Delta m^{2}$. The power of the Planck CMB constraint is robust to the choice of effective mass definition used in the cosmology model, giving similar results from the thermal and Dodelson-Widrow mechanisms. However, if we allow the lepton asymmetry to be very large ($L=10^{-2}$), the Planck exclusion region is significantly reduced. We also show the oscillation experiment constraints in the cosmology parameter space, where the same effect is observed. In this parameter space the MINOS constraints rule out a larger fraction of the region allowed by the CMB.

1607.00032

1607

1607.05329_arXiv.txt

{ In this contribution we will discuss the non-linear effects in the baryon acoustic oscillations and present a systematic and controllable way to account for them within {\it time-sliced perturbation theory}. }

\label{intro} Baryon acoustic oscillations (BAO) are widely used to establish the distance-redshift relation enabling to constrain the expansion history and the composition of the Universe \cite{Eisenstein:1998tu,Eisenstein:2005su}. BAO measurements aim at providing (sub-) percent precision in the near future and thus it is imperative to provide the best possible theoretical control over the BAO in order to fully exploit the potential of future surveys. The BAO are most prominent in the 2-point correlation function of matter density in position space where they form a peak at $r_{BAO}\sim $110 Mpc$/h$ in comoving coordinates. It has been observed long ago that the shape of the BAO peak retrieved from N-body simulations differs significantly from the prediction of linear theory (see Fig.~\ref{fig1}, left panel). Even though $r_{BAO}$ is significantly larger than the characteristic scale of non-linear clustering $2\pi k_{NL}^{-1}\sim 20$ Mpc$/h$, the leading correction computed in Eulerian standard perturbation theory (SPT) failed to capture the behavior seen in N-body data (see Fig.~\ref{fig1}, left panel). This disagreement is caused by large-scale bulk flows whose interaction with short modes is amplified if the distribution of matter has a feature. At leading order the effect of bulk motions is to worsen the correlation between galaxies at separations of order $r_{BAO}$, which results in the suppression of the BAO peak. Many approaches have been put forward in order to account for bulk flows. From the numerical side one can mention BAO {\it reconstruction} \cite{Eisenstein:2006nk}, which undoes bulk motions directly in the data and yields a sharper BAO peak with the signal-to-noise ratio improved by a factor of 2. From the theoretical side among the most successful approaches we would like to mention renormalized perturbation theory \cite{CrSc1,Crocce:2007dt}, Lagrangian perturbation theory \cite{Matsubara:2007wj} and IR - resummed effective field theory of large scale structures \cite{Senatore:2014via,Baldauf:2015xfa}. In this short contribution we will present a new way to systematically describe the non-linear evolution of BAO in the framework of time-sliced perturbation theory (TSPT) \cite{Blas:2015qsi,Blas:2016sfa}. \begin{figure} \begin{center} \includegraphics[width=0.49\textwidth]{xiall} \includegraphics[width=0.49\textwidth]{pwiggly_rat} \end{center} \caption{\label{fig1} {\it Left panel}: Two-point matter correlation functions in position space at redshift zero, $\xi(r)\equiv \langle \delta(\textbf{r})\delta(0)\rangle$: N-body data from the Horizon Run 3 simulation \cite{Kim:2011ab} (red points), the prediction of linear theory (black dot-dashed line), and the 1-loop SPT result (blue solid line). {\it Right panel}: Ratio of oscillatory (wiggly) part $P_w$ of the linear power spectrum to the smooth part $P_s$. The $\Lambda$CDM cosmological parameters have been chosen as in \cite{Kim:2011ab}. } \end{figure}

\label{sec:conclus} In these notes we discussed the non-linear effects in the BAO and sketched the way how one can systematically take them into account within time-sliced perturbation theory. We outlined the physical picture behind the interactions with large scale bulk flows and argued the need for IR resummation if one works within the Eulerian framework. Then we made a short introduction into the TSPT formalism and discussed its key virtues relevant for IR resummation: a clear way to separate the perturbative expansion into the smooth and wiggly components, and manifest IR safety of the TSPT loop integrands. We introduced the power counting rules which were used to identify and resumm relevant sets of diagrams at leading and next-to-leading orders. Finally, we compared our results with N-body data and found good agreement within data errors. We point out that the TSPT framework can be easily extended to incorporate higher-order corrections due to non-linear clustering, as well as new physics, e.g. the effects of neutrino masses or primordial non-gaussianity. {\bf Acknowledgments} Author is grateful to D. Blas, M. Garny and S. Sibiryakov for their contribution to this work. The work is supported by the Swiss National Science Foundation.

1607.05329

1607

1607.00817_arXiv.txt

Future observations of 21~cm emission using HI intensity mapping will enable us to probe the large scale structure of the Universe over very large survey volumes within a reasonable observation time. We demonstrate that the three-dimensional information contained in such surveys will be an extremely powerful tool in searching for features that were imprinted in the primordial power spectrum and bispectrum during inflation. Here we focus on the ``resonant'' and ``step'' inflation models, and forecast the potential of upcoming 21~cm experiments to detect these inflationary features in the observable power- and bispectrum. We find that the full scale Tianlai experiment and the Square Kilometre Array (SKA) have the potential to improve on the sensitivity of current Cosmic Microwave Background (CMB) experiments by several orders of magnitude.

While generic slow-roll models of cosmic inflation predict a nearly scale-invariant power spectrum of primordial curvature perturbations, there exist also many theoretically motivated implementations of the inflationary mechanism that predict \emph{features}, i.e., significant local deviations from scale invariance~\cite{Chluba:2015bqa}. Power spectrum features are typically accompanied by a correlated, similarly strongly scale-dependent signal in higher-order spectra (e.g., \cite{2007JCAP...06..023C,2012JHEP...05..066A,2013JCAP...10..038B}), which in principle allows us to discriminate between different scenarios by combining power spectrum and bispectrum information~ \cite{2015PhRvD..91b3502F,2016PhRvD..93d3536M}. However, analyses of present cosmic microwave background (CMB) anisotropy data have not found evidence for such features in the power spectrum~\cite{2015arXiv150202114P} or bispectrum~\cite{2015arXiv150201592P,2016PhRvD..93d3536M,2015PhRvD..91l3506F,2015arXiv151208977A} with a statistical significance higher than 3$\sigma$, after accounting for the look-elsewhere effect. It is therefore worth enquiring whether other observables may be more suitable for the detection of such features. CMB lensing aside, the temperature and polarization maps of the CMB only provide us with 2-dimensional information about cosmic perturbations. This not only imposes a fundamental limit to how precisely we can predict the angular power spectra (i.e., cosmic variance), but also obscures features through the necessary geometrical projection effect. The large scale structure (LSS) of the Universe, on the other hand, is accessible to tomographic measurements, which retain the 3-dimensional information of the perturbations. For a sufficiently large survey volume, cosmic variance can be pushed beyond the CMB limit. Constraints on features models have previously been discussed in the context of using the galaxy power spectrum~\cite{2011PhRvD..84h3505C,2012JCAP...04..005H,2015PhRvD..91f4039H}. Here we investigate the potential of detecting inflationary features in primordial density perturbations using sky surveys with the redshifted 21 cm emission from neutral hydrogen (HI), especially the 21cm intensity mapping observations. In the intensity mapping mode of observation, individual galaxies or clusters are not resolved, only the total 21cm intensity of large cells which contains many galaxies are measured \cite{2008PhRvL.100i1303C}. What the intensity mapping survey loses in angular resolution it makes up for in survey speed, allowing us to potentially cover unprecedented survey volumes, and it has been shown to have exquisite sensitivity to various cosmological parameters \cite{2015ApJ...803...21B, 2015ApJ...798...40X}. A number of 21cm intensity mapping projects have been proposed, such as the single dish array feed BINGO (BAO from Integrated Neutral Gas Observations) project \cite{2013MNRAS.434.1239B}, and the cylinder arrays CHIME (Canadian Hydrogen Mapping Experiment) \cite{CHIME} and Tianlai (Chinese for ``heavenly sound") projects \cite{2012IJMPS..12..256C}. Intensity mapping survey is also being considered for the upcoming Square Kilometer Array (SKA) phase one mid-frequency dish array (SKA1-MID) \citep{2015MNRAS.450.2251Y}. Below we shall study the full scale Tianlai array and the SKA1-MID cases. We investigate two observables: the 21 cm \emph{power spectrum} and \emph{bispectrum} respectively, and focus on two models with oscillatory features: the \emph{resonant} model and the \emph{step} model.

Very recently the potential of greatly improving constraints on oscillatory features in power spectrum with future large scale structure observations was noted in Refs.~\cite{2016arXiv160509365C,2016arXiv160603747B}, which investigated the potential of Euclid and LSST galaxy power spectrum observations, and Ref.~\cite{2016arXiv160509364C}, which looked at future 21~cm measurement through the dark ages. Here we show that the upcoming 21 cm intensity mapping observations of the LSS in the post-reionization Universe {\it alone} could put extremely tight constraints on the feature models. While the exact limit derived from the observation may depend on the details of the survey, such as the redshift range, sky area, system temperature and total observation time, and the precision actually achieved may be somewhat lower than the forecast due to less-than-perfect foreground removal, these surveys would still make orders-of-magnitude improvements over the two-dimensional CMB measurements. Furthermore, we also considered the bispectrum measurements, which were not previously considered for galaxy surveys, and found that it could also provide constraints better than the CMB. In addition, the sensitivity may be further improved by combining the power spectrum and bispectrum measurements. \\\\

1607.00817

1607

1607.05745_arXiv.txt

A growing class of ultra-high energy neutrino observatories based on the Askaryan effect and Antarctic ice is able to search for Lorentz-invariance violation. The ARA, ARIANNA, ANITA, and EVA collaborations have the power to constrain the Standard-Model Extension by measuring the flux and energy distribution of neutrinos created through the GZK process. The future expansion of ARA, at the South Pole, pushes the discovery potential further.

Ultra-high energy neutrino (UHE-$\nu$) observations are a long-desired achievement in astroparticle physics. Clues about cosmic ray origins and potential electroweak interaction measurements from $10^{16}$-$10^{19}$ eV are contained within this flux.\cite{kotera} PeV-scale neutrino observations in IceCube\cite{ice,ice2} have made possible learning about UHE-$\nu$ physics from beyond the solar system. A UHE-$\nu$ could be produced via the GZK process, given the UHE-$p^{+}$ spectral cutoff at $10^{19.5}$ eV.\cite{TA} The next generation of UHE-$\nu$ detectors is designed around the Askaryan effect, which produces radiated radiofrequency power.\cite{zhs,rb,arz,jch} Antarctic ice provides a convenient medium for Askaryan radiation.\cite{icejch} The RICE collaboration\cite{rice} began the field, and efforts such as ANITA, ARA, ARIANNA, and proposed EVA\cite{anita,ara,arianna,eva} have made progress in developing sensitivity to UHE-$\nu$ fluxes. There is a connection between Lorentz-invariance violation (LIV) and UHE-$\nu$, through flux limits, via the Standard-Model Extension (SME).\cite{sme} The SME includes LIV terms of varying dimension, proportional to small coefficients. LIV in the neutrino sector could modify the UHE-$\nu$ spectrum at Earth by introducing vacuum energy loss.\cite{gorham} The UHE-$\nu$ detectors can place constraints on SME coefficients.

1607.05745

1607

1607.02353_arXiv.txt

We present the first spectropolarimetric observations of a hydrogen-free superluminous supernova at \mbox{$z=0.1136$}, namely \ae. The transient shows significant polarization at both the observed epochs: one 24 days before maximum light in the rest-frame, and the subsequent at 27 days after peak luminosity. Analysis of the Q-U plane suggests the presence of a dominant axis and no physical departure from the main axis at either epoch. The polarization spectrum along the dominant axis is characterized by a strong wavelength dependence and an increase in the signal from the first to the second epoch. We use a Monte Carlo code to demonstrate that these properties are consistent with a simple toy model that adopts an axi-symmetric ellipsoidal configuration for the ejecta. We find that the wavelength dependence of the polarisation is possibly due to a strong wavelength dependence in the line opacity, while the higher level of polarisation at the second epoch is a consequence of the increase in the asphericity of the inner layers of the ejecta or the fact that the photosphere recedes into less spherical layers. The geometry of the superluminous supernova results similar to those of stripped-envelope core collapse SNe connected with GRB, while the overall evolution of the ejecta shape could be consistent with a central engine.

\label{sec:intro} The last six years have seen the surprising discovery of new classes of intrinsically bright supernovae. They show absolute peak magnitudes of M$\sim-21$ \citep[e.g.][]{qu11,2011ApJ...743..114C,in14} and a tendency to occur in dwarf, metal poor galaxies \citep[e.g.][]{ch13,lu14,le15a,ch16}. They do not exhibit the typical narrow spectroscopic features of strongly interacting supernovae and they are usually referred as superluminous supernovae \citep[SLSNe; see the review of][]{gy12}. They can be divided in two main groups, hydrogen free - and hence labeled SLSNe I - and hydrogen rich - therefore called SLSNe II. The first class is better studied: they show blue continua at maximum light, a distinctive ``W'' feature due to O~{\sc ii} at early epochs and, at about 30 days after peak they are spectroscopically similar to normal or broad-lined type Ic SNe at peak luminosity \citep{pa10}; consequently, they are also usually labeled SLSNe Ic \citep[e.g.][]{in13}. Additionally, SLSNe Ic show different light curve behavior and, as a consequence, have been divided in the subgroup of fast evolving \citep[e.g. SN~2005ap, SN~2010gx;][]{pa10,qu11} and that of slow evolving \citep[e.g. SN~2007bi, PTF12dam;][]{gy09,ni13}. The SLSNe of type II are fewer in number and they also show blue continua at maximum light, together with broad hydrogen features. At about 20 days after peak they show some resemblances to normal type IIL, although by definition they are several magnitudes brighter \citep[see][for a review]{in16}. Despite the increasing number of objects found every year, the nature of their explosion and the progenitor scenario are still debated. The favored scenario for all the types is that of an explosion driven by a magnetar as a central engine, which deposits energy into the supernova ejecta and significantly enhances the luminosity \citep{wo10,kb10,de12}. However, alternative scenarios such as the accretion onto a central black hole \citep{dk13}, the interaction with a dense circumstellar medium \citep[CSM,][]{ch12} and a pair instability explosion are still feasible alternatives. The pair-instability mechanism is only physically plausible for some of the objects, in particular those with the slowest evolving lightcurve \citep[e.g. see][]{gy09,koz15} A powerful diagnostic to distinguish between scenarios is polarimetry since it can unveil information on the geometry of the explosion and hence increase our understanding of the transients. Imaging polarimetry of a fast evolving SLSNe Ic has been reported by \citet{le15} but no evidence of asymmetries was found. However, spectropolarimetry offers a more in depth analysis of the geometry of SN explosions \citep[see][for a review]{ww08}. Quantitative modelling of the lightcurves of SLSNe Ic indicates that the data are well matched by the explosion of massive progenitors with a central engine \citep[e.g. as in ][]{in13,ni13}. If this were true, we would expect to see an intrinsically asymmetric explosion characterized by a dominant polarization angle as observed for stripped envelope SNe \citep[e.g.][]{wa01,ma07,tan12}. Indeed many core collapse SNe show large-scale departures from axisymmetry \citep[e.g.][]{le06,ww08}. If a magnetar is the central engine that powers the extreme luminosity,, then the strong magnetic field (B$\sim10^{14}$G) and rapid rotation could lead to asymmetries \citep{chk16}. Such asymmetries are potentially stronger than in normal stripped envelope SNe. Detection of such signatures may suggest a core-collapse origin combined with an asymmetric magnetar energy injection process. On the other hand, alternative scenarios such as the interaction with a CSM would exhibit spectropolarimetric evolution similar to those of type IIn SNe \citep[e.g.][]{le00,wa01}. These signatures are different than that of stripped envelope SNe, since they typically show unpolarised broad lines and loops across spectral lines arising from the CSM \citep{ww08}. An axisymmetric ejecta could be the consequence of: an aspherical production of energy and momentum from an explosion due to magnetohydrodynamic jets \citep{ko99} or magnetoturbulence \citep[e.g.][]{mo15}, accretion flow around the central neutron star \citep[e.g.][]{chev89}, asymmetric neutrino emission \citep[e.g.][]{wha10,mu15} or a combination of these; or the fact that the material could be ejected in clumps. In this paper we present our spectropolarimetric observations, as well as the interpretation of the geometry of the brightest known superluminous supernova, namely \ae\/, which belongs to the group of slow evolving SLSNe Ic.

We have presented the first spectropolarimetric data for a SLSN. We have gathered two epochs, one pre-peak at -23.7d and the second 27.5 days after maximum light in the rest-frame. Our analysis of these data indicates: \begin{enumerate} \item the presence of a dominant axis and no physical departure from it; \item a strong wavelength dependence of the polarization spectra; and \item an increase in the mean degree of polarization from the first to the second epoch of observations \end{enumerate} We used our Monte Carlo toy-code to compare with the data in order to interpret the wavelength dependence and the time evolution of the polarization level. We calculated polarization spectra for prolate ellipsoidal geometries accounting for electron scattering and resonant line scattering. We were able to reproduce the wavelength dependence of the polarization by adopting a line opacity distribution for the inner ejecta that is rich in iron-group elements. This poses some challenges for the observed flux spectra, but further comprehensive modelling with full radiative transfer is required. We have also show that the evolution of the overall pseudo-continuum level can be replicated via two-zone aspherical ejecta models with an outer zone having A$_{\rm out}=0.88$ at both epochs and an inner zone showing an increase in the asymmetry from the first (A$_{\rm in}=0.88$) to the second (A$_{\rm in}=0.60$) epoch. These values are calculated for an equatorial viewing angle and thus should be considered as lower limits on the asphericities of the ejecta. Orientations away from the equatorial plane would lead to lower polarization signals and would then require smaller axis ratio values. In addition, we caution that the specific values derived here (A$_{\rm in}=0.88-0.60$ and A$_{\rm out}=0.88$) are model dependent. However, our findings qualitatively demonstrate that the \ae\/ spectropolarimetry can be reproduced by an ellipsoidal geometry - or alternatively a bipolar geometry - with an inner region that increases its asymmetry as time passes. The general trend for an increase in polarization with time (and the implied geometry) are reminiscent of those observed in stripped-envelope core-collapse SNe connected with $\gamma$-ray bursts, supporting the possible link between SLSNe Ic and ultra long $\gamma$-ray bursts. Among all the suggested scenarios for SLSNe I, these new data tend to favor a core-collapse explosion and a central inner engine as explanation for the inferred axisymmetric geometry, as well as the increase of the inner asymmetry. This could be achieved with an explosion of a rotating stellar progenitor. Currently, this scenario does not comfortably reproduce the wavelength dependence we observe, since it seems challenging to obtain enough line opacity. On the other hand, we did not explore other opacity options than a simple iron core and hence a more complete modelling is needed (including self-consistent spectra and polarimetry). Despite the uncertainties that still exist with our relatively simple modelling approach, the data should drive new initiatives in how to model these data with hydrodynamic simulations. These models will address questions such as the the geometry (bipolar or ellipsoidal shapes) and how an engine driven or interaction scenario could explain the wavelength dependence we observe (which we interpret as due to line opacity). We also need to know if other SLSNe I and broad line SLSNe II show similar properties. The next step is to collect further observations of the future nearby ($z<0.2$) SLSNe at least at two epochs, ideally one before and one post maximum light. Predicting polarization signatures for multi-dimensional hydrodynamic explosion models will then be required to better investigate the geometry of these explosions.

1607.02353

1607

1607.07740_arXiv.txt

Pulsars are the only cosmic radio sources known to be sufficiently compact to show diffractive interstellar scintillations. Images of the variance of radio signals in both time and frequency can be used to detect pulsars in large-scale continuum surveys using the next generation of synthesis radio telescopes. This technique allows a search over the full field of view while avoiding the need for expensive pixel-by-pixel high time resolution searches. We investigate the sensitivity of detecting pulsars in variance images. We show that variance images are most sensitive to pulsars whose scintillation time-scales and bandwidths are close to the subintegration time and channel bandwidth. Therefore, in order to maximise the detection of pulsars for a given radio continuum survey, it is essential to retain a high time and frequency resolution, allowing us to make variance images sensitive to pulsars with different scintillation properties. We demonstrate the technique with Murchision Widefield Array data and show that variance images can indeed lead to the detection of pulsars by distinguishing them from other radio sources.

While pulsars are primarily detected and observed with high time resolution in order to resolve their narrow pulses, the phase-averaged emissions of many pulsars can be detected in previous radio continuum surveys~\citep[e.g.,][]{kca+98,ht99,k00}. More importantly, continuum surveys are equally sensitive to all pulsars, not affected by the dispersion-measure (DM) smearing, scattering or orbital modulation of spin periods, and therefore allow us to search for extreme pulsars, such as sub-millisecond pulsars, pulsar-blackhole systems and pulsars in the Galactic centre. A number of attempts have been made to search for pulsars in radio continuum surveys~\citep[e.g.,][]{kcc+00,ckb00}. Although the majority of these attempts have been unsuccessful, the first ever millisecond pulsar discovered, B1937+21, was initially identified in radio continuum images as an unusual compact source with a steep spectrum~\citep{bkh+82}. Next-generation radio continuum surveys, such as the ASKAP-EMU (Australian SKA Pathfinder-Evolutionary Map of the Universe)~\citep{nha+11}, LOFAR-MSSS (LOFAR-Multifrequency Snapshot Sky Survey)~\citep{hpo+15} and MWATS (Murchison Widefield Array Transients Survey)~\citep{bck+13}, will map a large sky area at different radio frequencies with high sensitivities (e.g., $\sim$10\,$\mu$Jy for EMU at $\sim1.4$\,GHz). Such surveys will necessarily detect a large number of pulsars in the images, and enable us to carry out follow-up observations and efficient targeted searches for the periodic signals. As we move towards the Square Kilometre Array (SKA) era, searching for pulsars in continuum images will complement the conventional pulsar search, and make it possible to find extreme objects. The main challenge of detecting pulsars in continuum surveys or Stokes I images is to distinguish them from other unresolved point radio sources. Continuum surveys such as EMU will identify $\sim7\times10^{7}$ radio sources, while there are only $\sim1.2\times10^{5}$ potentially observable pulsars in our Galaxy~\citep[e.g.,][]{fk06}. Searching for pulsations from a large number of candidates will be very time-consuming, and therefore we need good criteria to select pulsar candidates. Although we know that pulsars have steep spectra and high fractions of linear and circular polarisation, these criteria are not exclusive as galaxies can also have steep spectra and be highly polarised. Also, as we average emission over the pulse phase, linear and circular polarisation of pulsars can be significantly lower in continuum surveys~\citep[e.g.,][]{dhm+15}. However, pulsars are the only known sources compact enough to show diffractive interstellar scintillations (DISS), which distinguishes them from other radio sources. DISS are observed as strong modulations of pulsar intensities caused by the scattering in the ionised interstellar medium (IISM). The time-scales of DISS are of order of minutes and frequency scales are of order of MHz at $\sim1$\,GHz~\citep[e.g.,][]{r90}. Only recently have we had enough bandwidth and frequency resolution to detect DISS and next-generation continuum surveys will make it possible to search for pulsars as point sources showing strong intensity scintillations. \citet{crd96} first suggested searching for pulsars in variance images and pointed out that the variance of pulsar signals can be introduced by pulse to pulse variability and both interplanetary and interstellar scintillations. However, they only focused on detecting pulsars with variance in time caused by pulse to pulse variabilities, which have variation time-scales of orders of milliseconds to seconds. In their work (using the Molonglo Observatory Synthesis Telescope at 843\,MHz with a bandwidth of 3\,MHz) they were unable to search for frequency variations because of their restricted bandwidth. In this paper, we will focus on the modulation of pulsar intensity in both time and frequency caused by DISS, and investigate detecting pulsar with DISS in variance images. In Section~\ref{diss}, we briefly review the basics of pulsar scintillation. In Section~\ref{statistics}, we generally investigate the false alarm and detection probabilities and the detection sensitivity as a function of scintillation time-scales and bandwidths with simulations. In Section~\ref{demo}, we demonstrate the technique with data taken with MWA. We discuss our results and conclude in Section~\ref{discussion}.

\label{discussion} We investigated the variation of pulsar intensity caused by DISS in variance images. Through simulations, we studied the sensitivity of variance images on detecting pulsars and compared it with Stokes I images. Using data taken with MWA, we demonstrated that variance images can lead to the detection of pulsars and distinguish pulsars from other radio sources. We conclude that DISS of pulsars provides us with a unique way to distinguish pulsars from other radio sources. With the variance imaging technique, we will be able to select the most promising pulsar candidates from large-scale continuum surveys and enhance the efficiency of following targeted search. Variance images are most sensitive to pulsars whose scintillation bandwidth and time-scales are close to the channel bandwidth and subintegration time. Therefore, for a given continuum survey with certain total bandwidth and integration time, in order to achieve the highest sensitivity and detect as many pulsars as possible, we will need to retain frequency and time resolution as high as possible, and construct a set of variance images with different channel bandwidth and subintegration time. Typically the time scale is not a problem but it is common for the channel bandwidth to limit the detection of scintillation. On the other hand, increasing time and frequency resolution will decrease the overall sensitivity of variance images because the noise level in each channel and subintegration increases. This indicates that, for a given total bandwidth and integration time, variance images will be relatively less sensitive to pulsars with small scintillation bandwidth and time-scales. The sensitivity maps presented in the paper and simulations we developed can be used to predict the number of pulsars detectable with variance images for future large-scale continuum surveys, e.g., MWATS, EMU and SKA. Taking the sensitivity of EMU ($\sim10$\,$\mu$Jy) as an example, assuming we have enough bandwidth and time and frequency resolution to detect pulsar scintillation, the sensitivity of variance images constructed with EMU will be $\sim60$ to 100\,$\mu$Jy, depending on the time and frequency resolution. To determine the number of pulsars that can be detected, pulsar Galactic and flux density distributions and their scintillation time-scales and bandwidths will have to be considered. We defer studies of pulsar population and prediction of pulsar detection with variance images to future work. Although variance images allow unique identifications of pulsars, for given false alarm and detection probabilities, the sensitivity of variance images is lower than that of Stokes I images. Therefore, very faint pulsars detectable in continuum surveys might not be identified in variance images even if they show strong scintillation. However, the diffractive scintillation features of pulsars can still be powerful criteria to distinguish them from other radio sources. Instead of making variance images, we could first identify point sources in continuum surveys and then use scintillation features to distinguish pulsars from other sources.

1607.07740

1607

1607.05059_arXiv.txt

We have obtained a firm detection of Cyg X-1 during its hard and intermediate spectral states in the energy range of 40~MeV--60 GeV based on observations by the \fermi\/ Large Area Telescope, confirming the independent results at $\geq$60MeV of a previous work. The detection significance is $\simeq\!8\sigma$ in the 0.1--10~GeV range. In the soft state, we have found only upper limits on the emission at energies $\gtrsim$0.1 MeV. However, we have found emission with a very soft spectrum in the 40--80~MeV range, not detected previously. This is likely to represent the high-energy cutoff of the high-energy power-law tail observed in the soft state. Similarly, we have detected a \g-ray soft excess in the hard state, which appears to be of similar origin. We have also confirmed the presence of an orbital modulation of the detected emission in the hard state, expected if the \g-rays are from Compton upscattering of stellar blackbody photons. However, the observed modulation is significantly weaker than that predicted if the blackbody upscattering were the dominant source of \g-rays. This argues for a significant contribution from \g-rays produced by the synchrotron-self-Compton process. We have found that such strong contribution is possible if the jet is strongly clumped. We reproduce the observed hard-state average broad-band spectrum using a self-consistent jet model, taking into account all the relevant emission processes, e$^\pm$ pair absorption, and clumping. This model also reproduces the amplitude of the observed orbital modulation.

\label{intro} Cyg X-1, an archetypical black-hole binary, shows two main spectral states, hard and soft. In the hard state, the main component of its X-ray spectrum appears to be thermal Comptonization in a plasma with the electron temperature of $k T_{\rm e}\sim 100$ keV, which features a sharp cutoff [in the $EF(E)$ representation] at energies $E\gtrsim 200$ keV. Beyond $\sim$1 MeV, there is a clear high-energy tail, measured up to $\sim$3 MeV (e.g., \citealt{mcconnell02}, hereafter M02; \citealt*{jrm12,zls12}). The origin of the photon tail may be Compton scattering by a power-law electron tail above the thermal electron distribution in the accretion flow (e.g., M02). In the soft state, there is a strong disc blackbody component in the X-ray spectrum, peaking at $\sim$1 keV, followed by a pronounced high-energy tail, measured up to $\simeq$10 MeV (M02). \citet*{mzc13}, hereafter MZC13, detected high-energy \g-ray emission from Cyg X-1 based on observations by the Large Area Telescope (LAT) on board of \fermi. The detection was at a $\simeq\! 4\sigma$ significance, which corresponds to the chance probability of the detection of $\simeq\! 6\times 10^{-5}$ if the noise distribution is Gaussian. Moreover, that emission was present only in the hard spectral state, while only upper limits were found in the soft state. This ruled out the observed source being an artefact of the background subtraction, thus providing a strong argument for the emission actually coming from Cyg X-1. After the release of the new and much improved LAT calibration, {\sc Pass 8}, in the summer of 2015, and given the significantly increased on-source time with respect to the data analysed in MZC13, we embarked on a new analysis, which results we present here. During our work, the work by \citet{zanin16} (hereafter Z16) appeared. Independently of our results, they have obtained the detection of \g-ray emission from Cyg X-1, only in the hard state, at a $\simeq\! 8\sigma$ level. They also found some evidence for orbital modulation of the flux, expected if the emission originates, at least partly, from Compton scattering of stellar blackbody photons \citep{jackson72}, which process we hereafter abbreviate as BBC. Here, we provide results of our analysis of the emission, which extends the analysis of Z16 by presenting the discovery of strong emission at the softest measured energies in both hard and soft states, quantifying the orbital modulation in the hard state, and developing theoretical models of the \g-ray emission and its modulation. Furthermore, we present 15 GHz monitoring data from the Arcminute Microkelvin Imager (AMI), covering the entire duration of the LAT observations analysed here. The part of these data after MJD 57211 has not been published before. In our theoretical modelling, we adopt the parameters of Cyg X-1 similar to those in MZC13. The orbital period is $P\simeq 5.6$ d, and we assume the black-hole mass of $M_{\rm X}\simeq 16\msun$, and the (relatively uncertain) mass of the donor as $M_*\simeq 27\msun$ \citep{cn09,orosz11,ziolkowski05,ziolkowski14}, which correspond to the separation between the components of $a\simeq 3.2\times 10^{12}$ cm. We assume the stellar effective temperature of $T_*\simeq 2.55\times 10^4$ K and the luminosity of $L_*\simeq 4.8\times 10^{38}$ erg s$^{-1}$, which are the lower limits in the analysis of \citet{ziolkowski14}. We adopt the jet inclination with respect to the binary plane of $i\simeq 30\degr$ \citep{orosz11, ziolkowski14}, and the distance of $D=1.86$ kpc \citep{reid11}. The projected opening angle of the steady jet, present in the hard state, was constrained by \citet{stirling01} to $\lesssim 2\degr$. Given $i\simeq 30\degr$, the actual opening angle, $\Theta_{\rm j}$, is the above times $\sin i$, leading to $\Theta_{\rm j}\lesssim 1\degr$. We adopt here $\Theta_{\rm j}\simeq 0.5\degr$. The jet velocity is constrained by the lack of the counterjet to $\beta_{\rm j}\gtrsim 0.8$ \citep*{stirling01,zpr16}, and we assume here that lower limit. \begin{table*} \begin{center} \caption{The new LAT sources in the ROI. The tentative identifications are based on the SIMBAD database. The source {\tt n5} does not appear in Z16. } \begin{tabular}{cccccc} \hline Number& TS(3--30~GeV) & RA, Dec & Name in Z16 & Tentative ID & Source type\\ \hline {\tt n1} & 31 & 302.42, 35.68 & J2009+35 & HD 191612 & spectroscopic binary (O8), \\ &&&&&X-ray source (\xmm, {\it Rosat})\\ {\tt n2} & 18 & 301.24, 34.38 & J2005+34 & PSR J2004+3429 & pulsar\\ {\tt n3} & 27 & 298.69, 33.45 & J1955+33 & PN K 3--49 & planetary nebula\\ {\tt n4} & 24 & 297.26, 34.20 & J1949+34 & V* V1449 Cyg,& dwarf nova, X-ray source\\ &&&&1RXS J194917.1+341042\\ {\tt n5} & 31 & 297.60, 34.95 & -- & HD 226099, & spectroscopic binary (G5), \\ &&&&1RXS J194932.7+350119 & X-ray source\\ \hline \end{tabular} \end{center} \label{new} \end{table*}

\label{conclusions} We have obtained measurements and upper limits of the flux from Cyg X-1 in the 0.04--200 GeV energy band based on observations of the \fermi/LAT. Our observational results confirm, and significantly extend, the independent results of \citet{zanin16}. We have detected a steady emission in the 0.04--60 GeV energy band in the hard and intermediate spectral state. The 0.1--10 GeV emission, with the $8\sigma$ significance, can be approximately described as a power law with $\Gamma\simeq 2.4\pm 0.2$. In the soft spectral state, we have found upper limits at $E\gtrsim 80$ MeV, but we have detected a steep soft spectrum in the 40--80 MeV range. The measured 40--80 MeV flux is larger in the soft state than in the hard state (in the pattern opposite to that at higher energies), which argues against its origin from the local background. We discuss the issues related to the analysis at the lowest energies in Appendix \ref{soft}. We have found that the detections at $E\lesssim$0.1 GeV are well explained by the high-energy tails of the emission of the accretion flow, in both the hard and the soft state. The used models were published in 2002 and 2009 based on data at $E<10$ MeV only, but they still predict well the respective present measurements. This agreement further supports the reality of the LAT detection at the lowest energies. The measured spectra are relatively steep, and correspond to the high-energy cutoffs of the tails caused by the e$^\pm$ absorption. We have also quantified the orbital modulation of the \g-ray flux. We have not found any statistically-significant dependence of the modulation strength on energy at $E\geq 0.1$ GeV. The peak of the modulation was found at the orbital phases between about $-0.2$ and 0. The observed modulation is significantly weaker than that predicted if the blackbody upscattering were the dominant source of \g-rays. This argues for a significant contribution from \g-rays produced by the synchrotron-self-Compton process. We have found that such strong contribution is possible if the jet is strongly clumped. We have reproduced the observed hard-state average radio and \g-ray spectrum at $E\gtrsim 0.1$ GeV using a self-consistent jet model, taking into account all the relevant emission processes, e$^\pm$ pair absorption, and clumping. This model also reproduces the amplitude of the observed orbital modulation.

1607.05059

1607

1607.00951_arXiv.txt

We study the interaction of the early spherical GC wind powered by Type II supernovae (SNe~II) with the surrounding ambient medium consisting of the gaseous disk of a star forming galaxy at redshift $z\simgt 2$. The bubble formed by the wind eventually breaks out of the disk, and most of the wind moves directly out of the galaxy and is definitively lost. The fraction of the wind moving nearly parallel to the galactic plane carves a hole in the disk which will contract after the end of the SN activity. During the interval of time between the end of the SN explosions and the ``closure" of the hole, very O-poor stars (the Extreme population) can form out of the super--AGB (asymptotic giant branch) ejecta collected in the GC center. Once the hole contracts, the AGB ejecta mix with the pristine gas, allowing the formation of stars with an oxygen abundance intermediate between that of the very O-poor stars and that of the pristine gas. We show that this mechanism may explain why Extreme populations are present only in massive clusters, and can also produce a correlation between the spread in helium and the cluster mass. Finally, we also explore the possibility that our proposed mechanism can be extended to the case of multiple populations showing bimodality in the iron content, with the presence of two populations characterized by a small difference in [Fe/H]. Such a result can be obtained taking into account the contribution of delayed SN II.

\label{sec:introduction} Many spectroscopic, spectrophotometric and photometric observational studies have revealed that globular clusters (GCs) host multiple stellar populations characterized by different chemical properties. While the first stellar generation (FG) has a chemical composition equal to that of the pristine gas, the second generation (SG) is depleted in C and O, enhanced in N and Na, and in some cases is also characterized by significant differences in helium; with few exceptions, the iron abundance is the same in all the stellar populations, indicating that the ejecta of the Type II supernovae (SNe~II) belonging to the FG has been lost by the GC before the formation of the stellar SG \citep[][and references therein]{piotto15} { which otherwise should be very likely Fe-enhanced, at odd with what observed in most of the clusters.} In order to explain the picture provided by observational studies, a number of theoretical frameworks differing in the FG sources of the processed material out of which the SG forms have ben proposed: fast-rotating massive stars \citep{krause13}, massive interacting binaries \citep {demink09,bastian13a}, asymptotic giant branch (AGB) and super-AGB stars \citep{der08,der10}. Despite the numerous theoretical efforts, many questions remain unanswered as none of the proposed scenarios is exempt from problems and some arbitrary assumptions.r All the proposed models require the processed ejecta of the FG sources to be mixed and diluted with gas with pristine chemical composition. In this paper we focus on the dynamics of the dilution in the context of the AGB scenario, which has been originally proposed by \cite{der08} to account for the three populations, characterized by three different helium abundances, discovered in NGC\,2808. According to this scenario, the GC is initially composed by FG stars with the chemical abundances of the pristine gas out of which they formed; subsequently, the SN II explosions clear the cluster of the remaining gas, and the super-AGB and AGB ejecta can collect in the cluster centre through a cooling flow; the stars created within this cooling flow form the SG population. In order to reproduce the observed Na-O anticorrelation it is however necessary to mix pristine gas with the AGB ejecta \citep[see][]{der10,der11,der12}. In this framework, the only stars forming from pure AGB ejecta would be the very O-poor stars classified as Extreme (E) by \citet{carretta10} while stars with less extreme O depletion (those classified as Intermediate (I) by \citet{carretta10} ) would form from a mix of ejecta and pristine gas. For clusters hosting E SG stars, dilution should start after the beginning of the formation of the SG while in clusters where the E population is absent SG formation should involve both AGB ejecta and pristine gas. \citet{Renzini15} have suggested that AGB models are far from having explored the entire parameter space, and a possible revision of the cross section of the reaction rate destroying sodium could lead to a reassessment of the need for the dilution. { However, as already argued in \citet{der11}, dilution with pristine gas is likely to remain as an essential ingredient also in the case a revision of AGB models were to lead to yields characterized by a Na-O anticorrelation.} Understanding the origin of this pristine gas and its dynamics during the early evolutionary stages of globular clusters and the phase of SG star formation is a key aspect in the study of multiple-population globular cluster formation. In this paper we present a model aimed at following the evolution of the gas within a globular cluster forming in the disk of a high-redshift galaxy. We are first guided by the idea of quantifying under which hypothesis the cluster gas may satisfy the simplest constraints for the formation of multiple populations, that is: $i$) initially the pristine gas must be completely removed, at least from the clusters displaying an extended Na-O anticorrelation (see above); $ii$) the cluster should be able to reaccrete the pristine gas about 20--30 Myr after the end of the SN II explosions\footnote{This time interval derives from the chemical models of \citet{der10,der12}.}; $iii$) the pristine gas should not be contaminated by the SN II ejecta because in most clusters the different stellar populations have the same iron abundances We find out that a model fulfilling these requirements is compatible with the formation of clusters in the disks of galaxies at $z>2$ \citep[see e.g.][and references therein]{kravtsov05,kruijssen15} and may also explain why extreme populations are present only in massive clusters. Interestingly, the model may be easily extended to the case of multiple populations showing bimodality in the iron content, with the presence of two (or more) populations differing in [Fe/H] by a generally small fraction. While the effects of the explosions of type II supernovae `isolated in time' cannot be distinguished from the iron content of the clusters, which will be scarcely increased \citep{dantona16}, we show that there may be cases in which delayed-SNII events may stop the cooling flow for a period of some decades of million years, but their energy output is not able to push the ejecta much beyond the cluster, allowing, at the end of this epoch, a new burst of star formation to occur in gas polluted by both supernovae ejecta and AGB ejecta.

\label{sec:conc} We proposed a scheme for the interaction of the GC wind powered by the SNe II with the surrounding ambient medium. This interaction provides a viable mechanism to supply pristine gas and dilutes AGB ejecta during the SG star formation phase in early Globular Clusters. We have shown that our proposed mechanism naturally fulfills all the requirements necessary to produce the observed Na-O anticorrelation in the context of the AGB model. Initially, the spherical SN wind drives the expansion of a bubble that, in most cases, breaks out of the gaseous disk where the GC is located. As a consequence, almost all of the SN ejecta is lost along the vertical direction, while the portion of the wind moving mainly parallel to the midplane continues to drive the growth of a hole through the gaseous disk. Some time after the SN activity, the hole starts to contract carrying the surrounding ambient gas, {\it with pristine abundances}, toward the cluster. The contraction takes a certain period of time during which the ejecta of the massive AGB stars collect within the GC centre and Extreme, O-poor, He-rich stars can form. Once the ambient medium reaches the cluster, it mixes with the AGB ejecta, and Intermediate stars form out of this mixture. The results of the mechanism presented here are consistent with some still unexplained observational results. In particular, in agreement with observations, only clusters with large ${\cal L} = L_{\rm w}/n_0$, and thus mainly clusters with large initial masses, can form an extended Na-O anticorrelation, while the lower values of ${\cal L}$ expected for clusters with lower masses do not allow the formation of O-poor, He-rich Extreme stars. In general, a correlation between the spread in helium and the cluster mass is an outcome of the proposed model. Moreover, the pristine gas involved in the dilution is not enriched in iron, giving rise to a SG with the same [Fe/H] of the FG stars, as often found in the observed GCs. We finally conjectured that the presented scenario may also explain the s-Fe-anomalous clusters where the stars are divided into two main groups with different iron abundances, each group exhibiting an independent O--Na anticorrelation, and with the iron--richer group showing an enhancement of s--process elements. In fact, assuming that few tens of delayed SN II start to explode a short while after the end of the single SNe II explosions, and supposing that such activity lasts for 20--40 Myr, a second bubble forms which, however, stalls before breaking out. The SN ejecta remain trapped into the bubble, and fall back into the cluster when the bubble contracts. This leads to the formation of Intermediate stars in the same way as the collapse of the first bubble did; but the matter forming this second burst of SG stars is enriched with the SN ejecta, giving rise to a bimodality in the iron distribution of the stellar populations.

1607.00951

1607

1607.06093_arXiv.txt

We present novel constraints on cosmic-ray propagation in the Galaxy using the recent precise measurements of proton and helium spectra from AMS-02, together with preliminary AMS-02 data on the antiproton over proton ratio. To explore efficiently the large (up to eleven-dimensional) parameter space we employ the nested-sampling algorithm as implemented in the \textsc{MultiNest} package, interfaced with the \textsc{Galprop} code to compute the model-predicted spectra. We use VOYAGER proton and helium data, sampling the local interstellar spectra, to constrain the solar modulation potential. We find that the turbulence of the Galactic magnetic field is well constrained, i.e., $\delta=0.30^{+0.03}_{-0.02}(stat)^{+0.10}_{-0.04}(sys)$, with uncertainties dominated by systematic effects. Systematic uncertainties are determined checking the robustness of the results to the minimum rigidity cut used to fit the data (from 1$\,$GV to 5$\,$GV), to the propagation scenario (convection vs no convection), and to the uncertainties in the knowledge of the antiproton production cross section. Convection and reacceleration are found to be degenerate and not well constrained singularly when using data above 5$\,$GV. Using data above 1$\,$GV reacceleration is required, $v_{\rm A}=25\pm2$km/s, although this value might be significantly affected by the low-energy systematic uncertainty in the solar modulation. In a forthcoming companion paper, we investigate the constraints imposed by AMS-02 measurements on lithium, boron, and carbon.

Theory} The propagation of CR can be described by the well-known diffusion equation \cite{StrongMoskalenko_CR_rewview_2007} for the particle density $\psi_i$ of species $i$ per volume and absolute value of momentum $p$ \begin{eqnarray} \label{eqn::PropagationEquation} \frac{\partial \psi_i (\bm{x}, p, t)}{\partial t} = q_i(\bm{x}, p) &+& \bm{\nabla} \cdot \left( D_{xx} \bm{\nabla} \psi_i - \bm{V} \psi_i \right) \nonumber \\ &+& \frac{\partial}{\partial p} p^2 D_{pp} \frac{\partial}{\partial p} \frac{1}{p^2} \psi_i - \frac{\partial}{\partial p} \left( \frac{\diff p}{\diff t} \psi_i - \frac{p}{3} (\bm{\nabla \cdot V}) \psi_i \right) - \frac{1}{\tau_{f,i}} \psi_i - \frac{1}{\tau_{r,i}} \psi_i. \end{eqnarray} The various terms describe (i) spatial diffusion, usually assumed to be homogeneous and isotropic and thus described by the momentum-dependent diffusion coefficient $D_{xx}(p)$, (ii) convective winds, described by their velocity $\bm{V}(\bm{x})$, (iii) diffusive reacceleration, parametrized as a diffusion in momentum space with coefficient $D_{pp}(p)$, (iv) continuous energy losses through the coefficient $dp/dt=\sum_k dp_k/dt$ which sums over all the various processes, $dp_k/dt$, through which the particles lose energy, (v) adiabatic energy losses, present if $\bm{V}(\bm{x})$ has a nonzero divergence, and finally, catastrophic losses by (vi) decay or (vii) fragmentation, with decay and interaction times $\tau_r$ and $\tau_f$, respectively. The equation is typically solved assuming a steady state regime, meaning that $\psi_i$ does not depend on time and so the term on left-hand side is zero. Diffusion is naturally expected to be an energy dependent process, with particles being less deflected by the magnetic fields with increasing energy, and thus diffusing faster. This process is usually modeled by a power law in rigidity $R=p/|Z|$ (\cite{Blandford:1987pw}): \begin{eqnarray} \label{eqn::DiffusionConstant} D_{xx} &= \beta D_{0} \left( \frac{R}{4 \, \mathrm{GV}} \right)^{\delta}, \end{eqnarray} where $\delta$ is the index of the power-law, $D_0$ the overall normalization, and $\beta=v/c$ the velocity of the CRs; we set the normalization scale at 4$\,$GV. The constant for diffusive reacceleration $D_{pp}$ is usually related to the spatial diffusion $D_{xx}$ and to the velocity $v_\mathrm{A}$ of Alfven magnetic waves \cite{Ginzburg:1990sk,1994ApJ...431..705S} as \begin{eqnarray} \label{eqn::DiffusivReaccelerationConstant} D_{pp} = \frac{4 \left(p \, v_\mathrm{A} \right)^2 }{3(2-\delta)(2+\delta)(4-\delta)\, \delta \, D_{xx}}. \end{eqnarray} The amount of reacceleration is thus described in terms of the parameter $v_\mathrm{A}$. Finally, convective winds are assumed to be constant and orthogonal to the Galactic plane $\bm{V}(\bm{x})= {\rm sign}(z)\, v_{0,c} $. We note that, in principle, this parametrization implies an unphysical discontinuity at $z=0$. A smooth transition in the thin halo containing the sources (with size $\sim $0.2 kpc) would be more realistic. Nonetheless, since this parametrization has been widely employed in past works, we use it for the sake of comparison. The source term $q_i(\bm{x}, p)$ of primary CR is assumed to factorize into a species dependent normalization $q_{0,i}$, a space-depend part $q_{r,z}$ (where $r=\sqrt{x^2+y^2}$ and $z$ are Galactocentric cylindrical coordinates), and a rigidity dependent part $q_R$: \begin{eqnarray} \label{eqn::SourceTerm_1} q_i(\bm{x}, p) = q_i(r, z, R) = q_{0,i} \ q_{r,z}(r,z) \, q_R(R). \end{eqnarray} We model the rigidity dependence as double broken power law with smooth transitions \begin{eqnarray} \label{eqn::SourceTerm_2} q_R(R) &=& \left( \frac{R}{R_0} \right)^{-\gamma_1} \left( \frac{R_0^{\frac{1}{s}}+R^ {\frac{1}{s}}} {2(R_0)^{\frac{1}{s}} } \right)^{-s (\gamma_2-\gamma_1)} \left( \frac{R_1^{\frac{1}{s_1}}+R^ {\frac{1}{s_1}}} { R_1^{\frac{1}{s_1}} } \right)^{-s_1(\gamma_3 - \gamma_2)}, \end{eqnarray} where $R_0$, $R_1$ are the two break positions, $s$, $s_1$ the smoothing factors, and $\gamma_i$ ($i=1,2,3$) the slopes in the various rigidity ranges in between the breaks. The normalization is such that $q_R(R) =1$ at $R=R_0$. Typically, only one break has been considered in the literature, with value of the order $\sim 10\,$GV \cite{Trotta_CR_Propagation_2011}, or none\footnote{ In \cite{Putze_MCMC_CR_LeakyBox_2009,Putze_MCMC_CR_BoverC_2010} a source term $q \propto \beta^{-1} R^{-\gamma}$ is considered, which implies a break in momentum at a rigidity $\sim m/Z$, with an upward steepening of 1 in the slope.} \cite{Bernardo_Unified_CR_interpretation_2010}. On the other hand the recent discovery of a break at around $300\,$GV in the proton and helium spectra first by PAMELA \cite{Adriani_PAMELA_pHe_2011} and then by AMS-02 \cite{Aguilar_AMS_Proton_2015,Aguilar_AMS_Helium_2015} makes it necessary to introduce a second break for a proper description of the data. This was, indeed, considered, for example, in \cite{Johannesson_CR_Propagation_2016,Evoli:2015vaa}. We further introduced in \eqnref{SourceTerm_2}, as a novel feature with respect to previous studies, the parameters $s_i$ to explore the possibility of a smooth transition between the various regimes, as opposed to a sharp one. We mention here that an alternative possibility would be to model the break as a break in the diffusion rather than the injection spectrum. This has the same effect for the primaries' spectra but leads to different results for secondaries. The secondaries' injection spectra would reflect the break from the primary spectra, but the amount of the break would increase during the propagation, with the result that the break is expected to be twice as large as the one of primaries. Nonetheless, for antiprotons this effect would start to be significant only at very large energies (above few hundreds GV), which are not yet well measured by AMS-02, and thus the two scenarios are equivalent. The effect could be, instead, important for lithium or boron AMS-02 measurements, which extend to larger energies with respect to antiprotons. The spatial dependence, i.e., the source distribution, is parametrized as \begin{eqnarray} \label{eqn::SourceTerm_3} q_{r,z}(r, z) &= \left( \frac{r}{r_s} \right)^\alpha \exp \left( -\beta \frac{r-r_s}{r_s} \right) \exp \left( - \frac{|z| }{z_0} \right), \end{eqnarray} with parameters $\alpha = 0.5$, $\beta=1.0$, $r_s=8.5$ kpc, and $z_0=0.2$ kpc. For the analysis of $\gamma$ rays one usually uses source distribution inferred from pulsars \cite{Yusifov_Pulsar_Distribution_2004} or supernova remnants \cite{Case_SNR_Distribution_1998,Green:2015isa}. Typical parameter values in those cases are $\alpha \sim 1.6$, $\beta \sim 4$ with a flattening above $r\gtrsim10\,$kpc and a cutoff above $r\gtrsim30\,$kpc. We checked that changing the source distribution to those values has a negligible impact on the CR energy spectra after propagation. In the case of secondary CRs, as for antiprotons produced by primary CRs through spallation in the interstellar medium (ISM), the source term is given by the primaries themselves. More precisely the source term is the integral over the momentum-dependent production rate of the secondaries and the sum over the primary species $i$ and the ISM components $j$, \begin{eqnarray} \label{eqn::SourceTerm_pbar} q(\bm{x},p) = \sum\limits_{j=\mathrm{H,He}} n_j(\bm{x}) \sum\limits_{i=\mathrm{p,He}} \int \diff p_i \, \frac{\diff \sigma_{ij}(p, p_i)}{\diff p} \beta_i \, c\, \psi_i(\bm{x},p_i), \end{eqnarray} where $\sigma_{ij}$ is the antiproton production cross section by the species $i$ spallating over the ISM species $j$. The ISM is assumed to be composed of hydrogen and helium gas with fixed proportion 1:0.11. The abundance of secondaries is typically quite low with respect to the primaries, and this allows one to evaluate \eqnref{SourceTerm_pbar} with $\psi_j(\bm{x},p_i)$ calculated from \eqnref{PropagationEquation} neglecting in the first place the secondaries. \ac{Besides antiprotons, we will consider also secondary protons, i.e., primary protons that underwent inelastic scattering, losing a substantial fraction of their energy, and thus reappearing at low energies. We will also take into account tertiary antiprotons produced by the spallation of the secondary antiprotons during propagation. Secondary protons and tertiary antiprotons are described with the same formalism. Their source term can be calculated analogously to \eqnref{SourceTerm_pbar} but replacing $\psi_i(\bm{x},p_i)$ with the density of primary protons in the first case, and secondary antiprotons in the second case, and using the associated production cross section. The latter is approximated as the total inelastic non annihilating cross section of the incoming proton or antiproton times the energy distribution of the scattered particle, approximated as $1/E_{\rm kin}$. For more details see Ref.~\cite{Moskalenko:2001ya}.} To numerically solve the propagation equation \eqnref{PropagationEquation} and to derive the secondaries' and tertiaries' abundances we use the \textsc{Galprop} code\footnote{http://galprop.stanford.edu/} \cite{Strong:1998fr,Strong:2015zva}. We use version $r2766$\footnote{https://sourceforge.net/projects/galprop/} as basis, and we implement some custom modifications, such as the possibility to use species-dependent injection spectra, which is not allowed by default in \textsc{Galprop}. Furthermore, we allow for a smoothing of the originally simple broken power law as discussed above. The propagation equation \eqnref{PropagationEquation} is solved on a grid in the energy dimension and in the two spatial dimensions $r$ and $z$, assuming cylindrical symmetry of our Galaxy. The radial boundary of the Galaxy is fixed to $20\,$kpc, while the half-height $z_h$ is a free parameter. The radial and $z$ grid steps are chosen as $\Delta r=1\,$kpc, and $\Delta z = 0.2\,$kpc. The grid in kinetic energy per nucleon is logarithmic between $1$ and $10^7\,$MeV with a step factor of $1.4$. Free escape boundary conditions are used, imposing $\psi_i$ equal to zero outside the region sampled by the grid. We tested also more accurate choices for the above settings and found the results stable against the changes. \ac{Note also that we consider propagation of nuclei only up to $Z$=2, i.e., in practice, in \textsc{Galprop} we propagate $p$, $\bar{p}$, $^2$H, $^3$He, and $^4$He species plus the secondary protons and the tertiaries antiprotons. This also means that we neglect possible contributions from the fragmentation of $Z>$2 nuclei, which should be a good approximation since their fluxes are much lower than the $p$ and He fluxes. Nonetheless, in the specific case of our best-fit propagation scenario (see below), we verified explicitly that including nuclei with $Z>$2 in the calculation changes the spectra of He (i.e., $^3$He + $^4$He) only by few percent and protons (i.e., $p$ + $^2$H) by less than 1\%. This is also confirmed by the study in ref.\cite{Coste:2011jc}, where it is also shown that the $Z>$2 nuclei contribution to He is few \% (although the contribution to $^2$H, $^3$He can be, instead, up to 20-30\%).}

Summary and Conclusion} We have presented new constraints on the propagation of Galactic CRs from an (up to) 11-dimensional parameter fit to the latest AMS-02 spectra for $p$, He, and $\bar{p}/p$. Solar modulation is treated within the force-field approximation, but the modulation potential is constrained with a novel approach, fitting the unmodulated CR $p$ and He spectra to recently available low-energy data from VOYAGER, collected after the probe left the heliosphere and thus sampling the local interstellar CR flux. The VOYAGER data and the unmodulated spectra are fitted jointly to the AMS-02 data and the modulated spectra. As a first attempt, we try to fit the data with a universal injection spectrum for $p$ and He. We find that a universal injection is possible when fitting only $p$ and He data. In this case, the observed difference in $p$ and He slopes of about $\sim 0.1$ can be explained by a significant production of secondary $p$ so that the total primary plus secondary $p$ spectrum is steepened by the required 0.1 value in the slope. However, this requires a quite low value of the spectral index of diffusion $\delta\sim0.15$, and implies a large production of $\bar{p}$ which significantly overpredicts the observations. This scenario is, thus, in the final instance, not viable. For the main results we thus perform a fit leaving individual spectral freedom to $p$ and He. With this additional freedom a good fit to $p$, He, $\bar{p}/p$ spectra is achieved. The main result is a tight constraint on $\delta=0.30^{+0.03}_{-0.02}(stat)^{+0.10}_{-0.04}(sys)$, where the error is dominated by systematic uncertainties rather than statistical ones. The robustness of this result has been cross-checked against various factors, like the uncertainties in the solar modulation, the choice of the diffusion model framework, i.e., if convection is allowed or not, and the systematic uncertainties in the $\bar{p}$ production cross section. Since solar modulation is most important at low energies, its effect was studied using different cuts (1 and 5$\,$GV) on the AMS-02 data. The effect of uncertainties in the $\bar{p}$ production cross section was, instead, tested comparing the results of the fit when different available determinations of the cross section are used. Both of these effects have an order $10-20\%$ impact on the value of $\delta$, while the most important effect is the inclusion of convection in the model, which shifts the value of $\delta$ from $\sim 0.3$ to $\sim 0.4$. For the other propagation parameters the results are less definitive. The height of the Galactic halo and the normalization of diffusion present a well-known degeneracy, which, not surprisingly, cannot be resolved. In this respect, more precise ``CR-clocks'' measurements, like the ratio $^9$Be/$^{10}$Be, which will be available in the future from AMS-02, are necessary. Regarding convection and reacceleration, the fit above 5$\,$GV prefers large convection velocities of $v_{0,c} \agt 50\,$km/s and Alfven velocities of $v_\mathrm{A} \alt 25\,$km/s, with large parameter errors coming from a degeneracy between convection and reacceleration. The fit with data down to 1$\,$GV breaks this degeneracy and gives a well definite reacceleration of $v_\mathrm{A}= 25\pm2\,$km/s and preference for lower values of $v_{0,c} \alt 50\,$km/s. It remains, however, unclear how robust this determination of $v_\mathrm{A}$ is, since it relies on data below 5$\,$GV which are significantly affected by solar modulation. Finally, we find that a fit without convection is nonetheless possible, providing a good fit to the $p$, He and $\bar{p}/p$ data, and giving a similar value of $v_\mathrm{A}$. A comparison of these results from the constraints imposed from the AMS-02 observations of lithium, boron, and carbon will be presented in a forthcoming companion paper.

1607.06093

1607

1607.03952_arXiv.txt

We construct a cosmological scalar-tensor-theory model in which the Brans-Dicke type scalar $\Phi$ enters the effective (Jordan-frame) Hubble rate as a simple modification of the Hubble rate of the $\Lambda$CDM model. This allows us to quantify differences between the background dynamics of scalar-tensor theories and general relativity (GR) in a transparent and observationally testable manner in terms of one single parameter. Problems of the mapping of the scalar-field degrees of freedom on an effective fluid description in a GR context are discused. Data from supernovae, the differential age of old galaxies and baryon acoustic oscillations are shown to strongly limit potential deviations from the standard model.

\label{Introduction} In scalar-tensor theories the gravitational interaction is mediated both by a metric tensor and a scalar field. The interest in this type of theories of gravity is connected with the expectation that the observed late-time accelerated expansion of the Universe may be understood without a dark-energy (DE) component \cite{carroll,nojiri1,gannouji1}. Instead, it is the modified (compared with Einstein's theory) geometrical sector which is supposed to provide the desired dynamics\cite{copeland,torres,lobo,caldkam}. This may be seen as a geometrization of DE. Different aspects of scalar-tensor theories in general or subclasses of them have been investigated in \cite{catena,farao,CAPOZ,dolgov,chiba,duma,sofarao,soti,brookfield,abean,scalten,clifton,chibayam,joyce}. Scalar-tensor theories are formulated either in the Einstein frame or in the Jordan frame. Both frames are related by a conformal transformation. While matter and scalar field energies are separately conserved in the Jordan frame, the dynamics of both components is coupled in the Einstein frame for any equation of state (EoS) different from that of radiation. Because of the complex structure of scalar-tensor theories, simple solutions are difficult to obtain, even if the symmetries of the cosmological principle are imposed. Hence, in practice, the background expansion rate is usually obtained via numerical integration of the equations of motion. In general, the scalar-tensor-theory based cosmological dynamics may substantially differ from standard cosmology. Our focus here is on the simplest possible extension of the standard $\Lambda$CDM model that scalar-tensor theory can provide. In this minimalist approach we remain in the vicinity of the standard model at the present epoch and we aim to quantify the differences between scalar-tensor theory and general relativity (GR) by establishing a structure in which the scalar field $\Phi$ explicitly enters an analytic solution of the dynamics such that for $\Phi = 1$ the standard $\Lambda$CDM limit is recovered. To this purpose we construct a simple model which is analytically solved in the Einstein frame. With the help of a conformal transformation we then demonstrate how the field $\Phi$, which is given as a certain power of the scale factor, enters the (Jordan-frame) Hubble rate. Here we rely on an effective GR description of the Jordan-frame dynamics to determine the geometric equivalent of DE. In more detail, our starting point is a simple, analytically tractable expression for the coupling between nonrelativistic matter and the (Einstein frame) scalar field which modifies the standard decay of the matter energy density with the third power of the cosmic scale factor. This interaction-triggered deviation of the standard decay in the Einstein frame is modeled by a power-law behavior in terms of the Einstein-frame scale factor. Under this condition and if additionally an effective energy density and an effective pressure, linked by a constant ``bare" EoS parameter, are assumed, an explicit solution of the Einstein-frame dynamics is obtained with the mentioned power as an additional parameter. A straightforward conformal transformation then allows us to obtain the Hubble rate and the deceleration parameter in terms of this parameter in the Jordan frame as well. For the value zero of such new parameter, corresponding to $\Phi = 1$, both frames become indistinguishable and reproduce the dynamics of the standard $\Lambda$CDM model. Otherwise one has a variable $\Phi$ and the dynamics in both frames becomes different, deviating from that of the standard model. The analytic expression for the Hubble rate which explicitly clarifies the impact of the scalar field on the cosmological dynamics is the main achievement of this paper. We shall confront the deviations from the standard model with data from supernovae of type Ia (SNIa), the differential age of old galaxies that have evolved passively (using $H(z)$, where $H$ is the Hubble rate and $z$ is the redshift parameter) and baryon acoustic oscillations (BAO). The structure of the paper is as follows. In Sec.~\ref{Basic dynamics} we recall basic general relations for scalar-tensor theories and specify them to the homogeneous and isotropic case. In Sec.~\ref{two-component} we set up an effective two-component description in the Einstein frame, introduce our interaction model and find the Einstein-frame Hubble rate. The transformation to the Jordan frame is performed in Sec.~\ref{Jordan} where we also discuss the implications of a mapping of the scalar-field degrees of freedom on the effective fluid dynamics in a GR context. Section~\ref{observations} is devoted to a Bayesian statistical analysis on the basis of observational data of SNIa, $H(z)$ and BAO. Finally, in Sec.~\ref{conclusions} we summarize our results.

\label{conclusions} We have established a class of scalar-tensor-theory-based cosmological models which are simple extensions of the $\Lambda$CDM model. The background dynamics of this class has been discussed in detail. Our main result is an analytic expression (\ref{H2Phi}) for the Hubble rate which explicitly quantifies the difference to the standard model through a constant parameter $m$ which determines the dynamics of the scalar field $\Phi$. The solution for $\Phi$ is a consequence of the solution of the macroscopic fluid dynamics. It corresponds to a scalar-field dynamics (\ref{eqPhieff}) with an effective potential given by (\ref{Ueff}). We identified a geometric equivalent of the DE component of the standard model. The corresponding effective energy density may be positive or negative, including a transition between both regimes. The effective EoS parameter of this component diverges at the transition point but the overall dynamics is well behaved. A similar comment holds for the effective adiabatic sound speed square of the geometrical DE. Such behavior is unavoidable in any model in which the energy density of an DE equivalent and its time derivative change their signs during the cosmic evolution. This seems to limit the usefulness of an effective fluid description of parts of the geometrical sector and may cause computational problems. But since the overall dynamics and the dynamics of the matter component are smooth, the mentioned apparently unwanted features do not really jeopardize the model. They just demonstrate the fact that the fluid picture for parts of the geometry is a formal description which may well differ from that of a real fluid. Our tests of the model parameter $m$ with data from SNIa, $H(z)$ and BAO constrain $m$ to values very close to the $\Lambda$CDM value $m=0$ with a slight preference for positive values. We expect the existence of the analytic background solution (\ref{H2Phi}) to be useful for future investigation of the perturbation dynamics of this scalar-tensor extension of the $\Lambda$CDM model. Even a very small non-vanishing value of $|m|$ will certainly modify the standard scenario of structure formation.\\ \ \\ \noindent {\bf Acknowledgement:} Financial support by CNPq, CAPES and FAPES is gratefully acknowledged.

1607.03952

1607

1607.01020_arXiv.txt

In most studies of dust in galaxies, dust is only detected from its emission to approximately the optical radius of the galaxy. By combining the signal of 110 spiral galaxies observed as part of the Herschel Reference Survey, we are able to improve our sensitivity by an order-of-magnitude over that for a single object. Here we report the direct detection of dust from its emission that extends out to at least twice the optical radius. We find that the distribution of dust is consistent with an exponential at all radii with a gradient of $\sim -1.7$\,dex\,$R_{25}^{-1}$. Our dust temperature declines linearly from $\sim$25\,K in the centre to 15\,K at $R_{25}$ from where it remains constant out to $\sim$2.0\,$R_{25}$. The surface-density of dust declines with radius at a similar rate to the surface-density of stars but more slowly than the surface-density of the star-formation rate. Studies based on dust extinction and reddening of high-redshift quasars have concluded that there are substantial amounts of dust in intergalactic space. By combining our results with the number counts and angular correlation function from the SDSS, we show that with Milky Way type dust we can explain the reddening of the quasars by the dust within galactic disks alone. Given the uncertainties in the properties of any intergalactic dust, we cannot rule out its existence, but our results show that statistical investigations of the dust in galactic halos that use the reddening of high-redshift objects must take account of the dust in galactic disks.

Recent studies of the extinction of quasars \citep{Menard2010} and the reddening of galaxies \citep{Peek2015} imply that approximately 50\% of the interstellar dust in the universe lies outside galactic disks. However, we still know remarkably little about how far out the dust in galactic disks extends. Spiral galaxies have huge disks of atomic hydrogen that extend much further than the standard optical radius \citep[][$R_{25}$, the radius where the optical brightness corresponds to a \textit{B}-band brightness of 25\,mag\,arcsec$^{-2}$]{Bigiel2012}, but the atomic hydrogen is unprocessed material that may have fallen on the galaxies from outside. Studies into the extent of molecular gas using the carbon monoxide (CO) line, even when combining the results from a very large number of galaxies, have only managed to trace the molecular gas out to 1.1\,$R_{25}$ \citep{Schruba2011}. There have been many attempts to detect dust around nearby galaxies by using both its absorption and reflection properties in the optical, and its emission properties in the far-infrared. Observations of edge-on galaxies have been used to look for dust above the plane of the galaxy. \citet{Hodges-Kluck2014} found that UV emission extends between 5--20\,kpc from the mid-plane of the galaxies in their sample, and was consistent with the emission being light from the disk scattered by dust above the disk. They found that the halo dust has an exponential form with scale heights $\sim 2.5-5$\,kpc, much larger than typical scale heights of stellar thick disks \citep[0.3--1.5\,kpc][]{Hodges-Kluck2014}. \citet{Bocchio2015} combined \Hersc\ and \Spitzer\ data for the edge-on galaxy NGC\,891, finding extended emission above the disk with a scale height of 1.44\,kpc. The extent of dust disks has been investigated by looking for the effects of dust extinction of background sources (e.g., quasars, background galaxies) through the outskirts of a nearby galaxy. \citet{Holwerda2009} have shown in a pair of occulting galaxies that dust extinction can be reliably detected out to 1.5\,$R_{25}$. Attempts to detect dust in the outer disk from its emission were undertaken with the early infrared space telescopes, \IRAS\ and \ISO. However, they were limited by sensitivity, wavelength coverage and resolution. \citet{Nelson1998} obtained a 2$\sigma$ detection at 1.5\,$R_{25}$ using \IRAS\ by a stacking analysis, and \citet{Alton1998} used \ISO\ observations of a sample of eight nearby galaxies to show that the dust has a larger exponential scale-length than the optical emission. Recently \Spitzer\ and \Hersc\ data have been used to create radial profiles of the dust for the nearby galaxies in the SINGS/KINGFISH sample \citep{Munoz-Mateos2009b, Hunt2015}. The KINGFISH radial profiles presented in \citet{Hunt2015} typically extend to between 1.0--1.5\,$R_{25}$ at 250\micron, across a variety of morphological types. In this paper, we aim to go beyond the typical detection radius of dust for an individual galaxy, by combining the signal from a large sample of galaxies. Since the dust in the outskirts of galaxies is likely to be very cool ($\leq 15$\,K) with low surface brightness, the \textit{Herschel Space Observatory} \citep{Pilbratt2010} is the only telescope with enough sensitivity to have the potential to directly detect this emission. We have selected galaxies from the Herschel Reference Survey (HRS) \citep{Boselli2010}, the largest targeted survey of nearby galaxies with \Hersc. These galaxies all have either spiral or irregular morphology and lie face-on, giving us the best possible physical resolution. In Section~\ref{sec:data} we present the data used for this analysis. Section~\ref{sec:Method} describes our method for creating the averaged radial profiles of the dust and other galaxy components. Section~\ref{sec:results} presents our results, and the conclusions are presented in Section~\ref{sec:conc}.

\label{sec:conc} In this paper we present the results of combining images of a large sample of nearby galaxies made from observations with the \textit{Herschel Space Observatory}. By combining the images, we determine how the average surface-density of dust depends on radius. We find the following results: \begin{enumerate} \item We detect the emission of dust at $>$\,3$\sigma$ threshold out to a radius of 2.0\,$R_{25}$ at 250\micron\ and out to a radius of 1.75\,$R_{25}$ at 350 and 500\micron, by combining images of 117 galaxies from the HRS. When we restrict the sample to the 45 galaxies with $R_{25} > $1.5\arcmin\ we can trace dust at 250\micron\ out to a radius of 1.75\,$R_{25}$ at $>$\,3$\sigma$.\\ \item We fit modified blackbody SEDs to the SPIRE and PACS (100--500\micron) radial profiles. We find the dust temperature declines from $\sim$25\,K in the centre of the galaxy to $\sim$15\,K at 1.0\,$R_{25}$ where it remains constant out to 1.75\,$R_{25}$.\\ \item We find the radial dust-mass distribution in a galaxy is consistent with an exponential profile out to a radius of 1.8\,$R_{25}$. The gradient of the surface-density profile is -1.7\,dex\,$R^{-1}_{25}$.\\ \item We created profiles of the stellar mass (traced by 3.4 or 3.6\micron) and SFR (UV \& Infrared) using images from \Spitzer, GALEX and WISE. We found the stellar gradient is in good agreement with the dust gradient with the dust gradient between 1\% and 10\% larger (in dex\,$R^{-1}_{25}$) depending on the dataset used. The SFR profile is significantly steeper than that of the dust with a gradient of $\sim$-2.6\,dex\,$R^{-1}_{25}$.\\ \item Our results show a much steeper gradient in the surface density of dust compared to metals (based on typical metallicity gradients for nearby galaxies), similar to the results for individual galaxies \citep[e.g.,][]{Mattsson2012b}. In closed-box chemical evolution models this is a signature that dust originates from grain growth in the ISM.\\ \item We use our results for the radial distribution of dust in galaxies to see if the large extended halos of dust around galaxies as described by \citet{Menard2010} could have a significant contamination caused by the clustering of galaxies and the dust emission within individual galaxies. By using the number counts and angular correlation function from SDSS, we show that, assuming Milky Way type dust, the dust within galaxy disks might explain the reddening results. Because of the large uncertainties in our analysis and the even larger uncertainty of whether the dust in intergalactic space is like dust in galactic disks, we cannot rule out the the existence of intergalactic dust. We propose a simple method of testing whether there is actually intergalactic dust. \end{enumerate} Our results from a statistical stacking analysis with images from the \textit{Herschel Space Observatory} represent the most sensitive study of the extended dust emission around galaxies that will be possible until a new submillimetre space mission. It is likely to be impossible to improve on our results with ground-based telescopes because of the fluctuating emission from the atmosphere, which makes it difficult to detect low surface-brightness emission from galaxies.

1607.01020

1607

1607.06436_arXiv.txt

Measurement of the brightness temperature of extended radio emission demands knowledge of the gain (or aperture efficiency) of the telescope and measurement of the polarized component of the emission requires correction for the conversion of unpolarized emission from sky and ground to apparently polarized signal. Radiation properties of the John A. Galt Telescope at the Dominion Radio Astrophysical Observatory were studied through analysis and measurement in order to provide absolute calibration of a survey of polarized emission from the entire northern sky from 1280 to 1750 MHz, and to understand the polarization performance of the telescope. Electromagnetic simulation packages CST and GRASP-10 were used to compute complete radiation patterns of the telescope in all Stokes parameters, and thereby to establish gain and aperture efficiency. Aperture efficiency was also evaluated using geometrical optics and ray tracing analysis and was measured based on the known flux density of Cyg A. Measured aperture efficiency varied smoothly with frequency between values of 0.49 and 0.54; GRASP-10 yielded values 6.5\% higher but with closely similar variation with frequency. Overall error across the frequency band is 3\%, but values at any two frequencies are relatively correct to $\sim$1\%. Dominant influences on aperture efficiency are the illumination taper of the feed radiation pattern and the shadowing of the reflector from the feed by the feed-support struts. A model of emission from the ground was developed based on measurements and on empirical data obtained from remote sensing of the Earth from satellite-borne telescopes. This model was convolved with the computed antenna response to estimate conversion of ground emission into spurious polarized signal. The computed spurious signal is comparable to measured values, but is not accurate enough to be used to correct observations. A simpler model, in which the ground is considered as an unpolarized emitter with a brightness temperature of $\sim$240 K, is shown to have useful accuracy when compared to measurements.

\label{intro} Full exploitation of a radio telescope requires detailed and accurate knowledge of its radiation properties. Modern tools of antenna engineering, electromagnetic simulators, offer the ability to analyze telescope performance. To what extent can they deliver a true picture of telescope characteristics? Can we use them to calculate telescope behavior to useful accuracy? In this paper we investigate these questions by analyzing the properties of a large radio telescope and comparing the computations with measurements. The telescope that we study is the John A. Galt Telescope at the Dominion Radio Astrophysical Observatory. We focus on two problems particularly relevant to mapping the extended emission from the Milky Way, which is partially linearly polarized at decimeter wavelengths. First, we need to know the gain (or, equivalently, the aperture efficiency) of the telescope so that we can report our observations in units of brightness temperature. Second, we need to know the polarization behavior of the telescope so that we can correct observations for instrumental polarization. The John A. Galt Telescope (we will refer to it as the Galt Telescope) is a 26 m axially symmetric paraboloidal reflector. The Galt Telescope has recently been used to map the polarized emission from the entire northern sky over the frequency range 1280 to 1750~MHz as part of the Global Magneto-Ionic Medium Survey (GMIMS -- \citealp{woll09}). The results reported here have been used in processing the data from that survey. The challenge in astronomical polarimetry is the accurate measurement of a small polarized signal embedded in a (usually) much larger randomly polarized signal.{\footnote{A randomly polarized signal is one whose time average has no net polarization} Instrumental polarization has a deleterious effect on polarimetry because it converts an unpolarized signal into an apparently polarized one. Radiation from the ground is unavoidable in most reflector telescopes, and this emission is similarly converted into an unwanted polarized signal, usually a substantial one. We examine the properties of ground emission and the telescope response to it. To attain our two goals we compute the total radiation pattern of the telescope, including its polarization response. We describe telescope response in terms of Stokes parameters, widely used in astronomy to characterize the polarization state of a signal \citep{tinb96,wils14} but less frequently used in antenna engineering. Stokes parameter $I$ is proportional to the total intensity of the signal, parameters $Q$ and $U$ together describe the state of linear polarization and parameter $V$ describes the state of circular polarization. For a linearly polarized signal the polarized intensity is ${\rm PI}=\sqrt{{Q^2}+{U^2}}$ and the polarization angle is ${\alpha}={0.5~{{\rm{tan}}^{-1}({U/Q})}}$. Positive $V$ corresponds to right-hand circular polarization (RHCP) and negative $V$ to left-hand circular polarization (LHCP) \citep{ieee79}. The terminology of antenna engineering is rife with terms that were developed while considering the antenna as a transmitter, for example the use of {\it{feed}} to describe the antenna placed at the focus of a reflector, or {\it{spillover}} to describe radiation from the feed that enters the far field without encountering the reflector. We calculate the radiation properties of the Galt Telescope as a transmitter, and can confidently use the results to understand its behavior as a receiver because reciprocity informs us that the behavior of an antenna as a receiver is completely described by its properties as a transmitter. Our calculations assume that the antenna is transmitting a signal at a single frequency, but we apply our results to receiving the wideband noise signals of radio astronomy. In describing radiation patterns we use the terms $E$ plane and $H$ plane. The $E$ plane is the plane which contains the axis of a linearly polarized feed (or antenna) and the electric vector of the excitation. The $H$ plane is orthogonal to the $E$ plane, and also contains the antenna axis.

\label{disc} We have used GRASP-10 to calculate the properties of a large radio telescope, and have compared the results of those calculations with measurements wherever possible. The GRASP-10 evaluation of aperture efficiency of the Galt Telescope has successfully predicted the frequency dependence of that parameter, but the measured values are approximately 6.5\% lower than the calculated value. This difference amounts to an error of 0.3~dB in the calculation of the gain, which is $\sim{50}$~dB. While this is good accuracy for an engineering tool, radio astronomy hopes to do better. We also used simpler tools, ray tracing and geometrical optics, to examine aperture efficiency, and demonstrated that they too can provide useful accuracy and can provide useful insights into performance. We decomposed aperture efficiency into constituent efficiencies amenable to calculation on the basis of geometrical optics assumptions. The dominant effects in determining aperture efficiency are the illumination efficiency and blockage by the feed-support struts of the spherical wave from the feed: together they limit overall aperture efficiency to a maximum value of about 0.65. The ray tracing code was of particular value in estimating spherical wave blockage because it was designed to handle the cigar-shape of the support legs. Spherical wave blockage is severe because the axes of the feed-support struts are only $34^{\circ}$ from the telescope axis, and a substantial part of the aperture is shadowed by the struts. The principal weakness of GRASP-10 in this application is its inability to model the struts precisely. GRASP-10 models the struts as straight metal cylinders, while in fact they are cigar-shaped fiber glass and two of them carry metal sheathed cables. We modelled the struts as metal cylinders of diameter 25.4~cm. If we had increased this by 3\%, a mere 8~mm, the gain would have decreased by about twice this, 6\%, bringing the calculation close to the measured result. We have investigated the properties of a radio telescope over a very wide band, 1250 to 1750 MHz. The work that we describe here has established the absolute calibration of a wideband dataset. Previous work at these frequencies has been able to depend on absolutely calibrated horn measurements (e.g. the survey of \citealp{reic82} traced its calibration to \citealp{webs74}) but such measurements are confined to the narrow frequency bands allocated to radio astronomy. We have established an absolute scale of brightness temperature for the GMIMS survey. The overall error in this scale is 3\% but the relative error between any two frequencies in the band is less than this, about 1\%. We have used GRASP-10 to investigate instrumental polarization in the far sidelobes. Signals from the ground (and to a lesser extent from the atmosphere) that are inherently unpolarized or slightly polarized can be converted to apparently strongly polarized signals. Our analysis has shown that the principal routes of entry of these spurious signals are the spillover lobes and the scatter cones that are generated by the feed-support struts. The most rapid changes in spurious polarization occur when the scatter cone sidelobes pass from the sky to the ground, or vice versa. For maximum stability of instrumental polarization, observations should be planned to avoid telescope pointings where the scatter cones come close to the horizon. Our predictions based on GRASP-10 agree in intensity with the measured results, but there are differences in detail. Again, the less than perfect representation in GRASP-10 of the feed-support struts is a problem. We also have the fact that our ground model is very simple; it does not take into account surrounding buildings, some of which are quite close to the telescope. The ground is not a perfect absorber; it is a lossy dielectric. As a consequence its reflection and emission properties are polarization dependent. We have presented an analysis based on empirical data on ground polarization from studies using satellite-borne microwave radiometers. We conclude that the inherent polarization of ground emission is low at the angles that matter in the spillover lobes of the telescope. In fact our prediction of spurious polarization based on a polarized ground is not very different from our prediction assuming an unpolarized ground, and in a comparison with the measured data it is hard to say which route gives the better prediction. We conclude that the best estimate of ground brightness temperature at $\sim$1.5~GHz is about 240~K independent of angle of incidence. \citet{kalb10} derive a very similar value at 1.42~GHz. This conclusion does not conflict with observations that show that the Moon and planets are polarized at these frequencies (e.g. \citealp{zhan12,heil63}). A telescope observing the Moon or the planets samples only one angle of incidence using its main beam. In our case, however, the sidelobes of the Galt Telescope are sampling many different angles of incidence through many sidelobes. All these contributions are averaged and the ground appears to be almost unpolarized as seen by the sidelobes.

1607.06436

1607

1607.05954_arXiv.txt

{} {We present an innovative artificial neural network (ANN) architecture, called Generative ANN (GANN), that computes the forward model, that is it learns the function that relates the unknown outputs (stellar atmospheric parameters, in this case) to the given inputs (spectra). Such a model can be integrated in a Bayesian framework to estimate the posterior distribution of the outputs.} {The architecture of the GANN follows the same scheme as a normal ANN, but with the inputs and outputs inverted. We train the network with the set of atmospheric parameters ($T_{eff}$, $log\:g$, $\lbrack Fe/H \rbrack$ and $\lbrack \alpha/Fe \rbrack$), obtaining the stellar spectra for such inputs. The residuals between the spectra in the grid and the estimated spectra are minimized using a validation dataset to keep solutions as general as possible.} {The performance of both conventional ANNs and GANNs to estimate the stellar parameters as a function of the star brightness is presented and compared for different Galactic populations. GANNs provide significantly improved parameterizations for early and intermediate spectral types with rich and intermediate metallicities. The behaviour of both algorithms is very similar for our sample of late-type stars, obtaining residuals in the derivation of $\lbrack Fe/H \rbrack$ and $\lbrack \alpha/Fe \rbrack$ below $0.1$ dex for stars with Gaia magnitude $G_{rvs}<12$, which accounts for a number in the order of four million stars to be observed by the Radial Velocity Spectrograph of the Gaia satellite.} {Uncertainty estimation of computed astrophysical parameters is crucial for the validation of the parameterization itself and for the subsequent exploitation by the astronomical community. GANNs produce not only the parameters for a given spectrum, but a goodness-of-fit between the observed spectrum and the predicted one for a given set of parameters. Moreover, they allow us to obtain the full posterior distribution over the astrophysical parameters space once a noise model is assumed. This can be used for novelty detection and quality assessment.}

} The first automatic systems for spectral classification were developed in the 80s and based on two paradigms: expert systems and pattern recognition. Knowledge-based systems were created by defining a set of rules/models that relate, for the case of astronomical objects, spectral indexes (absorption line equivalent widths, colour indexes, etc.) with a spectral class in the MK classification system \citep{1943assw.book.....M}. Examples of these types of systems can be found in \cite{1981AJ.....86.1360T} and \cite{1982BICDS..23...39M}. Systems based on pattern recognition rely on a distance function (cross-correlation, euclidean, chi-squared) that is minimized between the observed spectra and a set of templates, so that the observed spectrum receives the class of the closest template. In the 90s, machine learning methods, specially artificial neural networks (ANNs), began to be applied for MK classification (see the works from \citealt{1994MNRAS.269...97V} and \citealt{1995ApJ...446..300W}). The use of ANNs offered several advantages. Firstly, it is not necessary to explicitly define spectral indexes and models to obtain the classification. Furthermore, once the ANN has been trained, its application is really fast in comparison with distance minimization schemes. Finally, ANNs provide accurate results even when the signal to noise ratio (S/N) of the spectra is very low. This property represents an important advantage in the analysis of extensive surveys, with a high percentage of low S/N data, as is the case with the Gaia survey. By the end of the 90s, a movement lead by researchers such as \cite{1997MNRAS.292..157B} and \cite{2001ApJ...562..528S}, changed the perspective of stellar spectral classification towards a process of astrophysical parameter (AP) estimation, which is a problem that resembles the nonlinear regression problem in statistics. ANNs are known to perform very well in nonlinear regression regimes, so they have remained in the state of the art for AP estimation. During the first decade of the 21st century, researchers such as \cite{2010PASP..122..608M} proposed the combination of ANNs and wavelets for improving the estimations as a function of the spectral S/N. One of the main criticisms of ANNs, as well of other machine learning schemes, is that they are incapable of providing an uncertainty measure on their solutions. Some authors have proposed schemes to provide confidence intervals in addition to the ANN outputs. To do this, it is necessary to take into account the different sources of error such as the training data density, target intrinsic noise, ANN bias, error in the observations acquisition, and the mismatch between training data and observations. Furthermore, such errors can be input dependent. For example, \cite{710658} presents an approximated Bayesian framework that computes the uncertainty of the trained weights, that can then be used to obtain the uncertainty predictions taking into account several sources of error that are input dependent. However, such a method requires the computation of the Hessian matrix, which is not feasible for large networks. The ANNs needed for AP estimation are usually very large because the network inputs are as many as the number of spectrum pixels. This number depends on the wavelength coverage and the spectral resolution, but usually is of the order of thousands. Other methods, such as bootstrapping, also require a large number of computations, which makes them unfeasible for large problems. Gaia, the astrometric cornerstone mission of the European Space Agency (ESA) was successfully launched and set into orbit in December 2013. In June 2014, it started its routine operations phase scanning the sky with the different instruments on board. Gaia was designed to measure positions, parallaxes and motions to the microarcsec level, thus providing the first highly accurate 6-D map of about a thousand million objects of the Milky Way. Extensive reviews of Gaia instruments modes of operation, the astrophysical main objectives and pre-launch expected scientific performance, can be found, for example, in \cite{2013hsa7.conf...82T}. A vast community of astronomers are looking forward to the delivery of the first non-biased survey of the entire sky down to magnitude 20. Moreover, the final catalogue, containing the observations and some basic data analysis, will be opened to the general astrophysical community as soon as it is be produced and validated. The definitive Gaia data release is expected in 2022-2023, with some intermediate public releases starting around mid-2016 with preliminary astrometry and integrated photometry. The Gaia Data Processing and Analysis Consortium (DPAC) is the scientific network devoted to processing and analysing the mission data. The coordination unit in charge of the overall classification of the bulk of observed astronomical sources by means of both supervised and unsupervised algorithms is known as CU8. This unit also aims to produce an outline of their main astrophysical parameters. The CU8 Astrophysical Parameters Inference System (APSIS, \citealp{2013A&A...559A..74B}) is subdivided into several working packages, with GSP-Spec (General Stellar Parameterizer - Spectroscopy) being the one devoted to the derivation of stellar atmospheric parameters from Gaia spectroscopic data. GSP-Spec will analyse the spectra obtained with Gaia Radial Velocity Spectrograph (RVS) instrument. Though its main purpose is to measure the radial velocity of stars in the near infrared CaII spectral region, it will also be used to estimate the main stellar APs: effective temperature ($T_{eff}$), logarithm of surface gravity ($log\:g$), abundance of metal elements with respect to hydrogen ($\lbrack Fe/H \rbrack$) and abundance of alpha elements with respect to iron ($\lbrack \alpha/Fe \rbrack$). The software package being developed by the GSP-Spec team is composed of several modules which address the problem of parameterization from different perspectives \citep{2006MNRAS.370..141R, 2010PASP..122..608M}, and has been recently described in \cite{2016A&A...585A..93R} (from now on, RB2016). This work focuses on developments carried out in the framework of one of these modules, called \textit{ANN}, that is based on the application of ANNs. During the commissioning stage of the mission (from February to June, 2014) unexpected problems were found that lead to a degradation of RVS limiting magnitude to a value close to $G_{rvs}=15.5$ mag \citep{2014EAS....67...69C}, that is around 1.5 mag brighter than expected. Figure \ref{fig:SNRvsGrvs} shows updated end-of-mission values for the $G_{rvs}$ versus S/N relationship for resolution element (3 pixels) that are based on simulations of RVS post-launch performance. The different algorithms for RVS stellar parameterization developed in the framework of Gaia DPAC need to be evaluated by the use of synthetic spectra at a variety of S/N values, which correspond to different magnitude levels. These values, then, already incorporate the revised performance figures. From RB2016 it was clear that ANNs give in general better results at very low S/N, this is one of the motivations of studying in detail such an approach for stellar parameterization, and also addressing the problem of ANN uncertainty estimations with Generative ANNs (GANNs). \begin{figure}[h] \centering \includegraphics[width=0.75\columnwidth]{images/SNRvsGrvs_comb.png} \caption{Signal to noise ratio as a function of the star magnitude $G_{rvs}$ for RVS post-launch configuration \citetext{D. Katz, 2015, private communication}} \label{fig:SNRvsGrvs} \end{figure} Uncertainty estimation of computed APs is crucial for the validation of the parameterization itself and for the exploitation of the results by the astronomical community. Therefore, some of the algorithms being developed in Gaia DPAC have addressed this problem. The idea is to change the perspective of the regression problem by learning the forward model (also called generative model) instead of the inverse model, which then allows the comparison between the observed spectrum and the spectrum estimated by the generative model. One proposed algorithm is Aeneas \citep{2012MNRAS.426.2463L}, that has been integrated in CU8 software chain APSIS. Aeneas defines a generative model that predicts the spectra from a set of APs by means of modelling with splines Gaia spectrophotometers data. Then, for a given spectrum, it finds the set of parameters that provide the estimated spectrum which maximizes the likelihood with respect to the observed one. To do so, it uses Markov chain Monte Carlo (MCMC, \citealt{10.2307/2346063}) algorithms to search for the best APs. The generative model is integrated in a Bayesian framework that enables the computation of the posterior distribution of the parameters given the observed spectrum. In this work, we also present a generative model but now based on neural networks, Generative ANNs, for AP estimation from Gaia RVS spectra. We discuss its performance in comparison with a classical ANN feed-forward algorithm. The remainder of this paper is organized as follows: in Section \ref{sect:simulations} we describe the library of synthetic spectra that is being used to test the algorithms, in Sections \ref{sect:anns_algo}, \ref{sect:ganns_algo} and \ref{sect:implem} we describe the ANN and GANN algorithms and their performances and in Section \ref{sect:results} we show the results obtained when it is applied to the Gaia RVS simulated data. Finally, in Section \ref{sect:discussion} we discuss the advantages and drawbacks of the proposed method.

} ANNs are a great tool that offer nonlinear regression capabilities to any degree of complexity. Furthermore, they can provide accurate predictions when new data is presented to them, since they can generalize their solutions. However, in principle, they are not able to give a measure of uncertainty over their predictions. Giving a measure of uncertainty over predictions is desirable in application domains where posterior inferences need to assess the quality of the predictions, specially when the behaviour of the system is not completely known. This is the case for data analysis coming from complex scientific missions such as the Gaia satellite. This work has presented a new architecture for ANNs, Generative ANNs (GANNs), that models the forward function instead of the inverse one. The advantage of forward modelling is that it estimates the actual observation, so that the fitness between the estimated and the actual observation can be assessed, which allows for novelty detection, model evaluation and active learning. Furthermore, these networks can be integrated in a Bayesian framework, which allows us to estimate the full posterior distribution over the parameters of interest, to perform model comparisons, and so on. However, GANNs require more computations, since the network needs to be evaluated iteratively before it reaches the best fit for the current observation. Some shortcuts that could reduce the number of required evaluations have been described, such as MCMC methods. In any case, the computation of AP uncertainties, taking into account all involved sources of errors, is possible with GANNs, while it is not feasible with other methods that use the Hessian matrix, at least without a significant implementation effort. The capability of both ANNs and GANNs to perform AP estimation was demonstrated by means of Gaia RVS spectra simulations, since they can efficiently contribute to the optimization of the parameterization. In most stellar types except metal poor stars, the parameterization accuracies are on the order of 0.1 dex in $\lbrack Fe/H \rbrack$ and $\lbrack \alpha/Fe \rbrack$ for stars with $G_{rvs}<12$, which accounts for a number in the order of four million stars. Internal errors here reported will need to be combined in the near future with the external uncertainties that will be obtained when real spectra from benchmark reference stars are analysed. GANNs give significantly better AP estimations than ANNs for A, B and early F stars, and only marginally improved for cooler stars. Nevertheless, its use in the regular Gaia pipeline of data analysis has been postponed to the next development cycle due to its high computational cost. The methodology presented here is not only valid for AP estimation, but is a general scheme that can be extrapolated to other application domains.

1607.05954

1607

1607.02059_arXiv.txt

The Hydrogen Intensity and Real-time Analysis eXperiment (HIRAX) is a new 400--800\,MHz radio interferometer under development for deployment in South Africa. HIRAX will comprise 1024 six\,meter parabolic dishes on a compact grid and will map most of the southern sky over the course of four years. HIRAX has two primary science goals: to constrain Dark Energy and measure structure at high redshift, and to study radio transients and pulsars. HIRAX will observe unresolved sources of neutral hydrogen via their redshifted 21-cm emission line (`hydrogen intensity mapping'). The resulting maps of large-scale structure at redshifts 0.8--2.5 will be used to measure Baryon Acoustic Oscillations (BAO). BAO are a preferential length scale in the matter distribution that can be used to characterize the expansion history of the Universe and thus understand the properties of Dark Energy. HIRAX will improve upon current BAO measurements from galaxy surveys by observing a larger cosmological volume (larger in both survey area and redshift range) and by measuring BAO at higher redshift when the expansion of the universe transitioned to Dark Energy domination. HIRAX will complement CHIME, a hydrogen intensity mapping experiment in the Northern Hemisphere, by completing the sky coverage in the same redshift range. HIRAX's location in the Southern Hemisphere also allows a variety of cross-correlation measurements with large-scale structure surveys at many wavelengths. Daily maps of a few thousand square degrees of the Southern Hemisphere, encompassing much of the Milky Way galaxy, will also open new opportunities for discovering and monitoring radio transients. The HIRAX correlator will have the ability to rapidly and efficiently detect transient events. This new data will shed light on the poorly understood nature of fast radio bursts (FRBs), enable pulsar monitoring to enhance long-wavelength gravitational wave searches, and provide a rich data set for new radio transient phenomena searches. This paper discusses the HIRAX instrument, science goals, and current status.

\label{sec:intro} % Measurements from Type Ia supernovae (SN1a)~\cite{SN1a}, Baryon Acoustic Oscillations (BAO)~\cite{2016MNRAS.457.1770C}, and the Cosmic Microwave Background (CMB)~\cite{2015arXiv150201589P} have shown that the energy density of the Universe is now dominated by Dark Energy, an unknown component causing the expansion rate of the Universe to increase. To better understand the nature of Dark Energy, we require measurements to a redshift of $z\sim2$, when Dark Energy began to influence the expansion rate. BAO provide a unique observational tool that can be extended to higher redshifts. BAO are a characteristic scale of 150\,Mpc in the matter power spectrum that is initially imprinted by primordial acoustic oscillations in the photon-–baryon fluid at decoupling. Because large-scale structure preferentially forms at that co-moving 150\,Mpc scale, measurements of the large-scale structure, and hence the BAO length scale, at various redshifts are sensitive tracers of the expansion rate of the Universe, allowing us to probe Dark Energy and its evolution. Recent galaxy surveys have already demonstrated percent-level constraints on Dark Energy at redshift $z\sim0.6$~\cite{2014MNRAS.441...24A} and have made a $5\sigma$ detection of high-redshift BAO using $\mathrm{Ly}\alpha$ forest measurements with quasars~\cite{2015A&A...574A..59D}.\newline A promising technique for measuring BAO at higher redshifts is 21-cm intensity mapping, whereby galaxies are observed in aggregate through low-resolution measurements of red-shifted 21-cm emission of neutral hydrogen. This method has two benefits: the 21\,cm emission line provides a natural redshift marker, and we can focus sensitivity on the large scales of interest to trace redshift evolution without resolving individual galaxies.~\cite{2015PhRvD..91h3514S}~\cite{2015aska.confE..19S}~\cite{2015ApJ...803...21B} Currently, 21\,cm intensity mapping has been measured only in the cross-correlation between radio surveys and galaxy surveys at redshift $z\sim0.8$: a $\sim$4$\sigma$ detection with the DEEP2 galaxy survey~\cite{2010arXiv1007.3709C}, and at higher significance with WiggleZ~\cite{2013ApJ...763L..20M}. Only upper bounds have been placed on the BAO contribution to the 21\,cm auto-correlation power spectrum.~\cite{2013MNRAS.434L..46S} \newline \begin{figure} [t] \begin{center} \begin{tabular}{c} % \includegraphics[height=7cm]{fbao_constraints_hirax.png} \end{tabular} \end{center} \caption[dish] { \label{fig:pspec} Expected baryon acoustic oscillations measurements with HIRAX, assuming a total sky area of 2,000 square degrees, 50\,K system noise, 1024 6\,m dishes, and 10,000 hours of integrated time, with Planck priors. We have divided by a fiducial model to make the power spectrum features more apparent. No systematics have been assumed. } \end{figure} The Hydrogen Intensity and Real-time Analysis eXperiment (HIRAX)\footnote{http:\//\//www.acru.ukzn.ac.za\//$\sim$hirax} is a radio interferometer being developed for observations in South Africa operating in the frequency range 400-800\,MHz. HIRAX will map large-scale structure in a redshift range of $0.8 < z < 2.5$ with 1024 six\,meter dishes. HIRAX will observe $\sim$15,000 square degrees in the Southern Hemisphere, complementing CHIME \cite{2014SPIE.9145E..22B} in the Northern Hemisphere. The resulting maps of large-scale structure will measure the first four peaks in the BAO matter power spectrum, reaching sample variance limits, as shown in Figure~\ref{fig:pspec}, and improve constraints on cosmological parameters relating to structure at late times (the amplitude of the matter power spectrum, $\sigma_{8}$; total baryon density, $\Omega_{b}$; total Dark Energy density $\Omega_{DE}$; and the power spectrum spectral index, $n_{s}$) to less than 1\% (Figure~\ref{fig:params}). Observations from experiments that span a wide range of redshifts, like HIRAX, allow us to measure both geometry and growth, provide a probe of dynamical Dark Energy, constrain modified gravity models, measure the isotropy of the Universe, and improve limits on deviations from Gaussianity of the initial density perturbations (See for example discussion in Ref.~\cite{2015aska.confE..19S}). Constraints on cosmological parameters require end-to-end simulations linking the underlying dark matter distribution with HI emission. In addition, these simulations will use combinations of analytical models, hydrodynamical N-body code (See Ref.~\cite{2016arXiv160401418D}) and treatment of instrumental effects and of polarized foregrounds (See Ref.~\cite{2014MNRAS.444.3183A}). \newline We will take advantage of the Southern location to overlap with many surveys at other wavelengths for cross-correlation science, including both ongoing surveys (ACTPol~\cite{2016arXiv160506569T}, SPTpol~\cite{2014SPIE.9153E..1PB}, SDSS BOSS~\cite{2013AJ....145...10D},DES~\cite{2016MNRAS.460.1270D}, HST~\cite{2014SPIE.9147E..5QW}) and future surveys (Advanced ACT\cite{2016JLTP..tmp..144H}, SPT-3G\cite{2014SPIE.9153E..1PB},DESI~\cite{2013arXiv1308.0847L}, LSST~\cite{2012arXiv1211.0310L}, Euclid~\cite{2012SPIE.8442E..0TL}, WFIRST~\cite{2013arXiv1305.5422S}). Cross-correlations with optical galaxy redshift surveys and cosmic shear surveys will provide measurements of the redshift-dependent neutral hydrogen fraction ($\Omega_{\mathrm{HI}}$) and bias ($\mathrm{b}_{\mathrm{HI}}$), both of which are poorly constrained at these redshifts, and probe the relationship between stars and gas in their dark matter halos. We can also use any combination of optical and neutral hydrogen galaxy distributions to remove cosmic variance, dramatically improving constraints in the future (See for example Refs. ~\cite{2009JCAP...10..007M}~\cite{2013MNRAS.432..318A}). These cross-correlations, including millimeter experiments like ACT and SPT, will also help identify and remove systematics between the different surveys and potentially allow intensity mapping surveys to better characterize the foregrounds and optimize their removal. Direct correlation with CMB surveys is challenging because of the loss of long-wavelength line-of-sight modes in the 21cm density field after foreground removal (see Section~\ref{sec:chall}), however, higher order correlations of the 21cm density field~\cite{moodley} or tidal field reconstruction\cite{2016PhRvD..93j3504Z} may provide an observable signal, allowing us to constrain the high-redshift matter power spectrum as well as $\Omega_{\mathrm{HI}}$ and $b_{\mathrm{HI}}$ independently of the optical galaxy bias. \begin{figure} [t] \begin{center} \begin{tabular}{c} % \includegraphics[height=9cm]{5pellipses_justHIRAX1.png} \end{tabular} \end{center} \caption[dish] { \label{fig:params} Forecast of cosmological parameter constraints for HIRAX assuming a total sky area of 2,000 square degrees, 50\,K system noise, 1024 6\,m dishes, and 10,000 hours of integrated time, with Planck priors.} \end{figure} Observations of the southern sky provide access to the Galactic plane, enabling a wide variety of transient measurements with HIRAX. A new window on the transient sky will be possible in the near future from upcoming wide-field imagers like LSST and gravitational wave detections of coalescing compact objects from LIGO~\cite{PhysRevLett.116.241103} and explosions. HIRAX will add radio transient monitoring to this suite of observations, which will open up the discovery space for transient phenomena and contribute to multi-messenger science, for example following up nearby explosive events found by LIGO~\cite{2013MNRAS.430.2121P}. In addition, Fast Radio Bursts (FRBs) are a source of enormous interest to the radio transient community because of their high dispersion measures and isotropic spatial distributions \cite{2016MNRAS.460.1054C}, indicating possible cosmological distances. Their origin is unknown and the subject of ongoing study (See for example Ref.~\cite{2015Natur.528..523M}). Projecting from current discovery rates of FRBs, HIRAX will find dozens per day (with significant uncertainty in the estimated number due to limited FRB statistics in the HIRAX band) and be able to measure properties associated with their spectra, pulse arrival times, and spatial distribution. In addition to FRBs, HIRAX can be used as a pulsar discovery engine and an efficient pulsar monitoring telescope. HIRAX pulsar searches have the potential to discover a few thousand new pulsars, while pulsar monitoring would provide important inputs to timing arrays used for the detection of long-wavelength gravitational waves, inaccessible to other probes.\cite{2010CQGra..27h4013H}. Finally, up-channelizing the data to a spectral resolution of $\sim$3 kHz (1.5\,km/s in the center of the band) will yield a spectral rms of $\sim$4.4\,mJy/beam. This will provide a competitive 21\,cm absorption line survey in a redshift range $1.36<\mathrm{z}<2.5$ which will not be covered by any of the upcoming radio surveys (MeerKAT\footnote{http:\//\//www.ska.ac.za\//meerkat}, ASKAP\cite{2012SPIE.8444E..2AS} and APERTIF\footnote{http:\//\//www.astron.nl\//general\//apertif\//apertif}). \newline In this paper we describe the HIRAX instrument (Section~\ref{sec:instru}) including its reflector design, analog chain, and digitization. Foreground removal and calibration challenges are described in Section~\ref{sec:chall}.

1607.02059

1607

1607.06485_arXiv.txt

The PRL Advanced Radial-velocity Abu-sky Search (PARAS) instrument is a fiber-fed stabilized high-resolution cross-dispersed echelle spectrograph, located on the 1.2 m telescope in Mt. Abu India. Designed for exoplanet detection, PARAS is capable of single-shot spectral coverage of 3800 - 9600 \AA, and currently achieving radial velocity (RV) precisions approaching $\sim$ 1 m s$^{-1}$ over several months using simultaneous ThAr calibration. As such, it is one of the few dedicated stabilized fiber-fed spectrographs on small (1-2 m) telescopes that are able to fill an important niche in RV follow-up and stellar characterization. The success of ground-based RV surveys is motivating the push into extreme precisions, with goals of $\sim$ 10 cm s$^{-1}$ in the optical and $< $1 m s$^{-1}$ in the near-infrared (NIR). Lessons from existing instruments like PARAS are invaluable in informing hardware design, providing pipeline prototypes, and guiding scientific surveys. Here we present our current precision estimates of PARAS based on observations of bright RV standard stars, and describe the evolution of the data reduction and RV analysis pipeline as instrument characterization progresses and we gather longer baselines of data. Secondly, we discuss how our experience with PARAS is a critical component in the development of future cutting edge instruments like (1) the Habitable Zone Planet Finder (HPF), a near-infrared spectrograph optimized to look for planets around M dwarfs, scheduled to be commissioned on the Hobby Eberly Telescope in 2017, and (2) the NEID optical spectrograph, designed in response to the NN-EXPLORE call for an extreme precision Doppler spectrometer (EPDS) for the WIYN telescope. In anticipation of instruments like TESS and GAIA, the ground-based RV support system is being reinforced. We emphasize that instruments like PARAS will play an intrinsic role in providing both complementary follow-up and battlefront experience for these next generation of precision velocimeters.

\label{sec:intro} % As we await the bounty of sources expected from space-based missions like TESS and GAIA, it is imperative that we tool up for the efficient and widespread ground-based follow-up of exoplanet candidates that will be required in complement. While confirmation of the most promising, and likely Earth analogous, candidates will be the prerogative of next-generation instruments like NEID and ESPRESSO, high-precision instruments on small (1-2~m) telescopes will be tasked with providing longer baseline observations on the rest. These instruments fill a niche that would be prohibitively expensive to fulfill with larger facilities, providing both high quality and quantities of data at desired cadence, limited only by the photon collecting power of small apertures. The data already gathered from such instruments (e.g. ELODIE, CORALIE, SOPHIE, CHIRON) are an important part of our extant canon of exoplanetary systems, and there is a growing necessity for more such workhorse RV instruments. Despite the recognition of this necessity, PARAS remains one of the few dedicated high-precision instruments of its caliber currently available to the community \cite{Fischer:2016}.

1607.06485

1607

1607.06957_arXiv.txt

Quasi-periodic pulsations (QPP) are often observed in X-ray emission from solar flares. To date, it is unclear what their physical origins are. Here, we present a multi-instrument investigation of the nature of QPP during the impulsive and decay phases of the X1.0 flare of 28 October 2013. We focus on the character of the fine structure pulsations evident in the soft X-ray time derivatives and compare this variability with structure across multiple wavelengths including hard X-ray and microwave emission. We find that during the impulsive phase of the flare, high correlations between pulsations in the thermal and non-thermal emissions are seen. A characteristic timescale of $\sim$20~s is observed in all channels and a second timescale of $\sim$55~s is observed in the non-thermal emissions. Soft X-ray pulsations are seen to persist into the decay phase of this flare, up to 20 minutes after the non-thermal emission has ceased. We find that these decay phase thermal pulsations have very small amplitude and show an increase in characteristic timescale from $\sim$40~s up to $\sim$70~s. We interpret the bursty nature of the co-existing multi-wavelength QPP during the impulsive phase in terms of episodic particle acceleration and plasma heating. The persistent thermal decay phase QPP are most likely connected with compressive MHD processes in the post-flare loops such as the fast sausage mode or the vertical kink mode.

\label{sec:intro} During a solar flare, the X-ray flux from the Sun can increase by several orders of magnitude and be accompanied by pulsations in the flare emission. These pulsations, known as quasi-periodic pulsations (QPP), are variations of flux as a function of time. Their nature has been examined in many previous studies of both solar \citep[e.g.][]{parks,kane,asai,fleisman, nak09, rez, li_ning} and stellar \citep[e.g.][]{balona, pugh} observations. In a typical event, the emission from a solar flare is seen to pulsate with a characteristic timescale ranging from $\leqslant$ 1~s up to several minutes. QPP are commonly observed during the impulsive phase of solar flares and have been reported in a wide range of wavelengths from radio and microwaves to hard X-rays (HXR) and $\gamma$-rays. Two main interpretations outlined in a recent review by \citet{nak09} have been pursued in order to understand QPP. These are that the observed flux variations are driven either by magnetohydrodynamic (MHD) wave behaviour in the corona, or by periodic or `bursty' energy releases from the coronal magnetic field. MHD oscillations are known to be supported in corona \citep{roberts82, roberts83}, and flare generated MHD oscillations have been observationally identified in coronal loops \citep{asch_99, nak_intro}. Various MHD wave modes can alter physical plasma parameters and produce the quasi-periodic behavior in flaring lightcurves \citep{nak09}. These wave modes can modulate the emission directly, affect the dynamics of charged particles, or periodically trigger magnetic reconnection \citep{nak_zim}. However it remains a challenge to interpret QPP purely in terms of linear MHD oscillation theory given the large modulation depths observed in lightcurves, along with the geometrical evolution that occurs during the impulsive phase of flares. \begin{figure*} \begin{center} \includegraphics[width=1.\linewidth]{Figure1.pdf} \caption{(a) Normalized light curves from different instruments for the flare of 2013 October 28. Detector NaI 6 was used for GBM. (b) Derivatives of the soft X-ray channels. The vertical red lines show the start and end of the impulsive phase and the dashed lines show the timing of the HXR pulses.} \label{fig:one} \end{center} \end{figure*} A second interpretation of QPP is that the pulsations are a direct result of `bursty' regimes of energy release, in particular magnetic reconnection. Early magnetotail studies suggested that magnetic reconnection can happen in an episodic fashion \citep{coppi, sch_coppi}, with subsequent observations supporting this view \citep{hones}. In coronal conditions, recent numerical models have shown that magnetic reconnection can occur in a repetitive regime \citep[e.g.][]{kliem, drake06, linton, guidoni}. Repeated episodes of magnetic reconnection itself could account for the modulation of emission in many different wavelengths. It could explain QPP observed in the non-thermal emission due to changing particle acceleration rates. Variations at other wavelengths such as soft X-ray (SXR) and EUV would then be explained by fluctuations in plasma heating. The majority of previous flare QPP studies have focused on pulsations observed in emission associated with non-thermal electrons such as HXR and microwave observations. This is due to the large modulation depths observed in this type of emission, especially in the impulsive phase. Recently, however, it has been shown that fine structure pulsations are also evident in the SXR emission \citep{dolla, simoes, dennis}. The nature of these pulsations in thermal emission remains to be studied in detail, and comparisons across multiple wavelengths are required to improve our understanding of the QPP phenomenon. In this letter, we investigate the nature of these X-ray pulsations in a multi-instrument analysis of a GOES X-class flare, paying particular attention to the fine structure observed during the impulsive and decay phases.

We have detected and analysed pulsations observed at multiple wavelengths during the X1.0 flare of 2013 October 28. Throughout the impulsive phase of the flare, highly correlated common features are observed at HXR, SXR, and microwave wavelengths with minimal time delay between peaks. Wavelet analysis of this impulsive interval shows broadband features in the wavelet power spectrum, with similar enhanced power in all channels. Characteristic peaks in the global spectrum at $\sim$20~s are detected in all wavebands with enhanced power also seen at around $\sim$55~s but only with significance above 99.7\% in the non-thermal emissions. These characteristic timescales are consistent with previous QPP investigations of different events. \citep[e.g.][]{Kupriyanova_2010, simoes, inglis_new} After the highly correlated impulsive phase, we find that emission in the non-thermal channels is no longer present. However, distinct pulsations in the high temperature plasma observed by GOES persist into the decay phase. The timescale of the pulsations are seen to increase from $\sim$40~s\ at the end of the impulsive phase at 02:00~UT to $\sim$70~s at 02:15~UT. These thermal pulsations could be a manifestation of continuing weak particle acceleration \citep[e.g.][]{maccombie}, or some other heating mechanism that persists into the decay phase of the flare. This would support the idea that continuous heating is required to describe the decay times observed in many flares which are longer than the estimated conduction and radiation cooling timescales \citep{ryan, cargill}. But what controls the timescale of the observed pulsations? The timescale of the QPP is consistent with expected characteristic timescales of MHD modes in the corona \citep{mcewan, pascoe07, pascoe09, macnamara}. However, given the complex evolution of geometrical plasma structures occurring during the impulsive phase, the identification of the specific MHD modes of oscillation producing the QPP is unlikely. The large modulation depths of the non-thermal emission during this time suggest that the pulsations are a result of episodic reconnection. The timescale would then be determined by either dynamic or periodic variations of the magnetic reconnection process such as multi-island reconnection in coronal current sheets \citep{drake06, guidoni}. During the decay phase however, the increase in timescale and small amplitude of the thermal SXR pulsations is consistent with MHD processes within the flaring site. Recent studies have attributed persistent SXR QPP to that of compressive MHD modes such as the global fast sausage mode \citep{tian} and vertical kink mode \citep{dennis}. Both of these compressive modes would result in the observed pulsations in the decay phase of this flare. The damped nature of the decay phase pulsations observed in Figure \ref{grad_deriv}(b) suggest that they are a result of the global sausage mode in the leaky regime. In the leaky regime, sausage modes are subject to damping and show a decaying oscillatory behaviour. The period of the sausage mode is determined by the ratio of the wavelength (twice the loop length $L$) to the external Alfv\'en speed: $P = 2L/V_{Ae}$ \citep{pascoe07}. This dependence decreases in the leaky regime, and in the long-wavelength limit, the period becomes independent of the length of the oscillating loop and is determined by the transverse travel time across the loop; $P^{leaky} \approx \pi a/V_{Ai}$, where $a$ is the loop width and $V_{Ai}$ is the internal Alfv\'en speed \citep{nak_12}. This may explain why the characteristic timescale stays constant at $\sim$70~s even when the loop length keeps increasing. Taking the period to be 70~s and $a$ in the range of $\sim$ 2-10'', this interpretation yields an estimation of $V_{Ai} \approx 65 - 350kms^{-1}$. This is considerably low for coronal loop Alfv\'en speeds \citep[e.g.][]{asch_vel}, suggesting that the loops are not sufficiently long enough to be in the long-wavelength limit and so this interpretation cannot fully explain our observations. However, given the low signal-to-noise in the late stage of the flare, it is difficult to determine the true nature of these pulsations. We cannot conclusively determine what mechanism generates these pulsations. However, the analysis of the SXR fine structure across multiple channels and its relation with other energies provides a new diagnostic tool. When correlations between different energies are high, especially SXR and HXR (as with most peaks in the impulsive phase of this flare), we can argue that the SXR emitting plasma is heated by electron beams at that time. When pulsations in SXR are seen to occur before HXR emission, and persist late into the decay phase after the HXR emission has stopped, some other heating mechanism is taking place. Thus, detailed comparisons between thermal and non-thermal structures can potentially help us understand the distribution of the different types of heating taking place during a flare. Additionally, it has now become apparent that fine structure pulsations observed in the SXR derivative are a common \citep{simoes}, and maybe even an intrinsic feature of flaring emission. Hence, further investigation into the structure evident in SXR lightcurves and multi-wavelength comparisons will provide better insight into the origins of the QPP phenomena.

1607.06957

1607

1607.00737.txt

In this article, we focus on a new scalar $\phi$ mediated scalar/vectorial WIMPs (weakly interacting massive particles) with $\phi$'s mass slightly below the WIMP mass. To explain the Galactic center 1 - 3 GeV gamma-ray excess, here we consider the case that a WIMP pair predominantly annihilates into an on-shell $\phi \phi$ pair with $\phi$ mainly decaying to $\tau \bar{\tau}$. The masses of WIMPs are in a range about 14 - 22 GeV, and the annihilations of WIMPs are phase space suppressed today. In this annihilation scheme, the couplings of the $\phi$ - standard model (SM) particles are almost arbitrary small, and the WIMP-nucleus spin-independent scattering can be tolerant by the present dark matter (DM) direct detections. A scalar mediator-Higgs field mixing is introduced, which is small and available. The lower limit on the couplings of the $\phi$-SM particles set by the thermal equilibrium in the early universe is derived, and this constraint is above the neutrino background for scalar DM in direct detections. The WIMPs may be detectable at the upgraded DM direct detection experiment in the next few years, and the exotic decay $h \rightarrow \phi \phi$, the production of $\phi$ may be observable at future high-luminosity $e^+ e^-$ collider.

The weakly interacting massive particle (WIMP) type dark matter (DM) attracts much attention in DM direct detections, and the cold DM relic density can be derived from thermally freeze-out WIMPs. Today, the compatible confident events are still absent in DM direct detection experiments, and the recent search results of CRESST-II \cite{Angloher:2015ewa}, CDMSlite \cite{Agnese:2015nto}, LUX \cite{Akerib:2015rjg} and XENON1T \cite{Aprile:2015uzo} set stringent constraints on the WIMP-nucleus spin-independent (SI) scattering. Even with these rigorous constraints, the case of the SI interaction being dominant in WIMP-nucleus scattering can still be allowed by the present direct detections, and a feasible scenario will be investigated in this work with the possible DM signatures from indirect detections. The cosmic ray observations, such as $\gamma$-rays, neutrinos, positrons, and antiprotons from DM dense regions, may indirectly reveal properties of WIMPs. The recent 1-3 GeV gamma-ray excess from the Galactic center may be due to WIMP annihilations, for WIMPs in a mass range about 35-50 GeV annihilating into $b\bar{b}$ with corresponding annihilation cross section $\sim (1-3)\times 10^{-26}$ cm$^3/$s \cite{Goodenough:2009gk,Hooper:2011ti,Abazajian:2014fta,Daylan:2014rsa,Calore:2014xka,Alves:2014yha,Zhou:2014lva}, or WIMPs in a mass range about 7 -11 GeV annihilating into $\tau \bar{\tau}$ with the annihilation cross section $\sim$ $0.5 \times 10^{-26}$ cm$^3/$s (20\% to $b\bar{b}$ also allowed) \cite{Hooper:2010mq,Hooper:2011ti,Abazajian:2014fta,Calore:2014xka,Daylan:2014rsa}. In this work, we focus on the latter case, that is, the main WIMP annihilation products in SM sector are $\tau \bar{\tau}$ pairs (see Refs. \cite{Lacroix:2014eea,Yu:2014mfa,Ibarra:2015fqa,Kim:2015fpa} for more discussions). Moreover, with a small number of visible matter in dwarf satellite galaxies, gamma rays from the DM-dominant dwarf galaxies provide significant information about WIMPs. The $\tau \bar{\tau}$ mode galactic center GeV gamma-ray excess can be compatible with the recent results from the new dwarf spheroidal galaxy observations \cite{Geringer-Sameth:2015lua,Drlica-Wagner:2015xua,Ackermann:2015zua,Li:2015kag}. New physics beyond the standard model (SM) is needed to yield the main product $\tau \bar{\tau}$ in SM sector in WIMP annihilations. The leptophilic WIMPs were discussed in the literature \cite{Baltz:2002we,Chen:2008dh,Pospelov:2008jd,Cholis:2008qq,Fox:2008kb,Cao:2009yy,Bi:2009uj,Ibarra:2009bm,Kopp:2009et,Cohen:2009fz}. Here we consider that a new scalar mediates the interactions between the SM charged leptons and scalar/vectorial WIMPs (the annihilation of fermionic WIMPs is p-wave suppressed today), and the new couplings of the mediator to leptons are proportional to the lepton masses. If the scalar mediator is lighter than the WIMP mass, the way of a WIMP pair annihilating into an on-shell mediator pair is allowed (see e.g. Refs. \cite{Martin:2014sxa,Abdullah:2014lla,Rajaraman:2015xka,Cline:2015qha} for more). In this case, the scalar mediator's couplings to SM particles can be almost arbitrarily small.\footnote{There is a lower bound about the couplings, which is from the thermal equilibrium in the early universe.} To fit the GeV gamma-ray excess and meanwhile evade present constraints from DM direct detections and collider experiments, we focus on the case that the mediator is lighter than the WIMP mass and the mediator-tau lepton coupling is much smaller than the mediator-WIMP coupling. Thus, the dominant annihilation mechanism of WIMPs is that a WIMP pair annihilates into an on-shell mediator pair which mainly decays to the heaviest leptons $\tau \bar{\tau}$. The case of $\tau \bar{\tau}$ mode dominant is naturally compatible with the antiproton spectrum observations from PAMELA \cite{Adriani:2010rc}, and is tolerant by the smooth positron spectrum of AMS-02 \cite{Bergstrom:2013jra,Hooper:2012gq,Ibarra:2013zia}. A small scalar mediator-Higgs field mixing is discussed, and the mixing is small enough to keep $\tau \bar{\tau}$ dominant in the scalar mediator decays. With a small mixing introduced, one prospect is that the WIMP-target nucleus SI scattering may be detectable at the upgraded DM direct detection experiment in the next few years, and another prospect is that the scalar mediator may be observable in the future high-luminosity $e^+ e^-$ experiment. In fact, the small mediator-Higgs mixing can play an important role to the thermal equilibrium between DM and SM sectors in the early universe. The reaction rates of SM particles $\rightarrow$ WIMPs should be larger than the expansion rate of the universe for some time in the early universe, and this sets a lower bound about the couplings of the scalar mediator to SM particles. The lower bound of the coupling gives a lower limit on the cross section of WIMP-target nucleus SI scattering. These will be explored in this paper. This article is organized as follows. After this introduction, the form of the interactions in new sector and the annihilation cross section of scalar/vectorial WIMPs are given in section II. Next we give a detailed analysis about scalar WIMPs in section III, including the constraints and the test at future experiment. In section IV, we give a brief discussion about the test of vectorial WIMPs. The conclusions and some discussions are given in the last section.

The scalar and vectorial WIMPs have been studied in this article, with a new scalar $\phi$ as the mediator and the mass of $\phi$ being slightly below the WIMP mass. The dominant annihilation products of WIMPs are on-shell $\phi \phi$ pairs with $\phi$ mainly decaying into $\tau \bar \tau$, and the WIMP annihilations are phase space suppressed today. For masses of WIMPs in a range about 14 - 22 GeV, the annihilation cross section $\langle\sigma_{ann} v_r \rangle_0 \sim$ $1 \times 10^{-26}$ cm$^3/$s today can be obtained to meet the Galactic center GeV gamma-ray excess. Due to the nearly arbitrary small couplings between $\phi$ and SM particles, the WIMP-target nucleus SI scattering can be tolerant by the present stringent constraints of DM direct detections. The upper limit of the $\phi$'s coupling to $\tau$ lepton is discussed with the constraints of WIMP annihilations, and the limit is tolerant by the muon $g - 2$ result. The scalar mediator-Higgs mixing angle $\theta$ should be small enough to keep $\tau \bar{\tau}$ dominant in the scalar $\phi$'s decay, and the upper limit of $\sin\theta$ from collider experiment is mild. The thermal equilibrium in the early universe sets an lower bound on the reaction rates of SM particles. Considering the $t \bar t$ contribution is dominant in SM particle reaction rates, we have derived an lower limit about the mixing angle $\theta$ and the coupling of the WIMP-$\phi$ trilinear term. For vectorial WIMPs, the reaction rate of $t \bar t \rightarrow$ WIMP pair is dramatically enhanced by the longitudinal polarization of vectorial WIMPs. For the scalar WIMP-nucleon SI elastic scattering of concern, we obtain that the bound from the thermal equilibrium sets a minimum scattering cross section above the neutrino background. The present parameter spaces are set by the XENON1T result and the bound from thermal equilibrium, and the allowed region of the elastic scattering cross section is derived, with $\sigma_{\rm {el}} \sim$ $10^{-48} - 10^{-46}$ $\rm{cm}^2$. The range of the scattering cross section can be examined at the future DM ultimate direct detection experiments, such as LUX-ZEPLIN (LZ) \cite{Akerib:2015cja}, XENONnT \cite{Aprile:2015uzo} and DARWIN \cite{Aalbers:2016jon}. Thus, for WIMPs of concern, the future DM direct detections can give an answer about whether the scalar WIMP candidates exist or not. For vectorial WIMPs, the bound from the thermal equilibrium is below the neutrino background in direct detections, and this type WIMPs cannot be ruled out by the future DM ultimate direct detections. The tests of the new sector at collider are as follows: i). The Higgs boson can decay into a WIMP pair, while this invisible decay is tiny and difficult to be explored at collider. ii). The decay channel $h \rightarrow \phi \phi$ may leave traces at the future $e^+ e^-$ collider with the precise Higgs decay measurement, e.g. via the Higgs-strahlung process at $\sqrt{s} = 250 $ GeV, and a high luminosity of 500 fb$^{-1} $ $-$ 5 ab$^{-1}$ is needed. iii). For $\sqrt{s} < $ 400 GeV and above the threshold, the production of $\phi$ is mainly via the $\phi$ strahlung mechanism at $e^+ e^-$ collider. The signature of $\phi$ with the $\theta$ value near the upper limit $\sin^2\theta = 10^{-3}$ may appear at the future $e^+ e^-$ collider, and it is better to test the non-standard model $\phi$-like particle at a low center of mass energy collider with a high luminosity, e.g. $\sqrt{s} \sim$ 120 - 150 GeV with the energy above the $Z \phi$ production threshold and the luminosity up to about 200 fb$^{-1}$. The future $e^+ e^-$ collider, such as the Circular Electron Positron Collider (CEPC) \cite{CEPC-Report}, the International Linear Collider (ILC) \cite{Baer:2013cma}, and FCC-ee(TLEP) \cite{Gomez-Ceballos:2013zzn}, may do the job to investigate the $h \rightarrow \phi \phi$ decay and the production of $\phi$. We look forward to the future tests of the WIMPs of concern via DM indirect detections, DM direct detections and the hunt at collider.

1607.00737

1607

1607.08384_arXiv.txt

{ We very briefly discuss proposed in the literature possible scenarios for intermediate mass black holes formation in globular clusters. We also discuss the results of the MOCCA simulations of about 2000 models (BigSurvey) regarding the distribution of events connected with electromagnetic and gravitational radiations, namely: mass transfer on IMBH, collisions and mergers with IMBH and mergers with IMBH due to gravitational radiation. The rates of these events are very small, so their observation is very improbable.}

Intermediate mass black holes (IMBH) are thought to be a missing link between stellar-mass black holes (BH) and supermassive BHs (SMBH). There are many theoretical arguments in favor of the formation of IMBHs in the centers of globular clusters (GCs) \citep[e.g.][and reference therein]{Lutzgendorfetal2013}. Observational confirmation of the existence of IMBHs would have an important impact on a number of open astrophysical problems related e.g. formation of SMBHs and their host galaxies, origins of ultraluminous X-ray sources in nearby galaxies, and detection of gravitational waves (GR). GCs are thought to be a natural site for IMBH formation. The proximity of GCs makes possible to directly observe kinematic and structural imprints of IMBHs. There are a lot of observations indirectly suggesting the presence of IMBHs in GCs in nearby galaxies or Milky Way. They are based on the detection of strong X-ray or radio emissions at of-centre positions in distant galaxies, not confirmed X-ray or/and radio emissions in some Galactic GCs or on kinematic and spatial structure of central parts of GCs. Up to now, there is no firm observational confirmation of IMBH presence in GCs.

The MOCCA code shows clearly its ability to model in a very efficient way a large number (thousands) of GC models. The observation of any events connected with accretion of matter on IMBHs or connected with mergers due to gravitational radiation is very improbable. We will have to have a lot of luck to see one of them. All the simulations data discussed here is a part of the BigSurvey database. This database can be freely accessed. If you are interested in using the data from the BigSurvey in your own research please send an email to Mirek Giersz.

1607.08384

1607

1607.03269_arXiv.txt

In this article, we compare a set of Wave Front Sensors (WFS) based on Fourier filtering technique. In particular, this study explores the "class of pyramidal WFS" defined as the 4 faces pyramid WFS, all its recent variations and also some new WFSs as the 3 faces pyramid WFS. Firstly, we describe such a sensors class thanks to the optical parameters of the Fourier filtering mask and the modulation parameters. Secondly, we use a unified formalism to create a set of performance criteria: size of the signal on the detector, efficiency of incoming flux, sensitivity, linear range and chromaticity. Finally, we show the influence of the previous optical and modulation parameters on these performance criteria. This exhaustive study allows to know how to optimize the sensor regarding to performance specifications. We show in particular that the number of faces has influence on the size of the signal but no influence on the sensitivity and linearity range. To modify these criteria, we show that the modulation radius and the apex angle are much more relevant. Moreover we observe that the time spent on edges or faces during a modulation cycle allows to adjust the trade-off between sensitivity and linearity range.

\label{sec:intro} By placing amplitude or phase masks in a focal plane, it is possible to filter the light from a pupil plane to another. Those masks are able, in particular, to transform incoming phase fluctuations into intensity variations on a detector. Such optical designs (see figure \ref{ff_bench}) are thus particularly relevant in the context of Wave Front Sensing, especially for the Adaptive Optics (AO). Moreover, Fourier based Wave Front Sensors (WFS) have many advantages compared to others Wave Front Sensors as for example the Shack-Hartmann in terms of, for instance, noise propagation or sampling flexibility. A general formalism about the Fourier based Wave Front Sensing has been recently developed in Fauvarque et al\cite{Fauvarque16}. Such a theoretical framework allows for instance to unify the Zernike Wave Front Sensor (WFS) introduced by Zernike\cite{zer1934} himself and the Ragazzoni's Pyramid WFS\cite{Rag96}. We choose in this article to explore, in the light of this formalism, the WFSs based on the Pyramid WFS principle, as for example the classical 4-faces modulated Pyramid, the 6- or 8-faces Pyramid introduced by Akondi et al.\cite{Akondi2014}, the Cone WFS (Vohnsen et al.\cite{Vohnsen2011}) or the Flattened Pyramid that we proposed in Fauvarque et al.\cite{Fauvarque2015}. The choice of studying this class of Fourier-based WFSs is due to the fact that the PWFS recently shows its great efficiency on sky on the Large Binocular Telescope (Esposito et al.\cite{LBT}), the Magellan Telescope (Close et al.\cite{MagAO}) and the Subaru Telescope (Jovanovic et al.\cite{Subaru}). It subsequently seems to be the most credible candidate for the next generations of Adaptive Optics, especially for the European Extremely Large Telescope. As a consequence, serious optimization works will have to be led in order to know which designs will be the most appropriate regarding to the Wave Front Sensing contexts. Finally, we mention some of the most enlightening theoretical works about this PWFS: Ragazzoni et al.\cite{ragazzoni1999sensitivity}, Esposito et al.\cite{Esposito1999}, Verinaud\cite{verinaud2004nature}, and Guyon\cite{guyon2005limits} which will serve as reference results along the article. In the first part of this paper, we describe the "Class of Pyramidal WFSs" which contains the modulated Pyramid WFS and all its variation when considering optical parameters of the mask and modulation settings as free parameters. The second part will define the unified performance criteria which will be used in the third part to compare all these Wave Front Sensors. In terms of optimization approach, the first part corresponds to the input parameters while the second part describes the output specifications. A final part will summarize the influence of each parameters on each performance criteria in order to create the best WFS regarding to an AO context. We recall here the general framework of the Fourier based Wave Front Sensing. The optical design is described in figure \ref{ff_bench}. \vspace{0.5cm} \begin{figure}[htbp] \begin{center} \includegraphics[width=16cm]{f1.eps} \end{center} \caption{Schematic view (in 1D) of a Fourier Filtering optical system. \label{ff_bench}} \end{figure} Such a device is a typical Fourier filtering system. The incoming perturbed electro-magnetic field $\psi_p$ is written: \begin{equation} \psi_p(\phi,n) = \sqrt{n}~\mathbb{I}_P~exp \left( \imath\phi \right) \end{equation} where $n$ is the spatial averaged flux, $\phi$ the perturbed phase at the analysis wavelength $\lambda_0$, and $\mathbb{I}_P$ is the indicative function of the entrance pupil. The Fourier mask which takes place on the focal point is considered as a pure transparent mask. Its transparency function may be written: \begin{equation} m = exp\left(\frac{2\imath\pi}{\lambda_0} OS\right) \end{equation} where $OS$ is the Optical shape of the mask. Via the Fraunhofer optical formalism, it is possible to write the intensity on the detector: \begin{equation} I(\phi,n) =|\psi_p(\phi,n) \star \mathcal{F}[m]|^2 \end{equation} where $\mathcal{F}[m]$ is the 2D Fourier transform of the mask. The phase seen by the WFS is the sum of the turbulent phase induced by the atmosphere and the static aberrations of the wave front sensing path. % Mathematically, we can split the incoming phase into two terms: \begin{equation} \phi = \phi_r + \phi_t \end{equation} where $\phi_t$ is the turbulent phase and $\phi_r$ the static reference phase. $\phi_r$ may also be seen as the operating point of the WFS. Thanks to a Taylor's development of the phase around the reference phase $\phi_r$: it is possible to get a phase power series of the intensity on the detector: \begin{equation} I(\phi,n)=n\Big(I_c+I_l(\phi_t)+I_q(\phi_t)+...\Big) \end{equation} The first term $I_c$ is the \emph{constant} intensity, it corresponds to the intensity on the detector when the phase equals to the reference phase, i.e., when the turbulent phase is null. The second term is the \emph{phase-linear} term. It corresponds to the perfectly linear dependence of the intensity regarding to the turbulent phase around the reference phase. The third term $I_q$ is the quadratic intensity, it corresponds to the first non-linear dependence of the intensity. The next terms are obviously non linear contributions as well. These equations may be easily generalized when a modulation device is working. The exact expressions of these intensities are given in Fauvarque et al.\cite{Fauvarque16} Since the signal on the detector is not linear with the phase, we have to process it. The easiest way to create a phase-linear quantity from $I(\phi,n)$ is to calculate the meta-intensity, called $mI$, via the following equation: \begin{equation} mI(\phi_t)= \frac{I(\phi_r+\phi_t,n)-I(\phi_r,n)}{n}\label{MI} \end{equation} The normalization by the factor $n$ allows to make $mI$ independent from the incoming flux. Retrieving $I(\phi_r,n)$ corresponds to a \emph{tare} operation. In practice, this \emph{return-to-reference} operation is done via a calibration path. This ensures that the meta-intensity equals to zero when there is no turbulent incoming phase. With such a definition, the meta-intensity $mI$ equals to the \emph{linear} intensity $I_l$ in the small phases approximation regime.

The numerous parameters of the sensors of the Pyramid class thus allows to consider the \textbf{optimization} of the element \emph{sensor} in an Adaptive Optics loop. The comparison of the different WFSs of this class, following unified performance criteria, shows that: -- The \textbf {number of faces} is essentially a geometric and technological parameter. Indeed, this parameter has no influence on the sensitivity and the linearity range. Consequently, its choice has to be done regarding sampling and manufacturing criteria. If the pupil images are completely separated, the 3-faces mask seems the best candidate since it minimizes the number of required pixels and is easy to manufacture. The Cone seems particularly relevant for the small apex angles. It is indeed perfectly chromatic due to the absence of edges and does not require a lot of pixels. -- The \textbf{tip/tilt modulation stage} provides two main advantages. It allows to retrieve a great part of the photon lost due to diffraction when the pupil images are separated. Moreover it improves drastically the linear range for the low spatial frequencies. Unfortunately a loss of sensitivity is associated to this gain. Simulations show that the shape of the modulation path (squared or circular) has no significant influence. Nevertheless, we observe that the sensitivity was linked to the time spent on the edges whereas the linearity range was correlated to the time spent on faces. Finally we note that modulation inevitably makes the WFS sensible to polychromatism. Without modulation all the sensors of the Pyramid class are rigorously achromatic. -- The \textbf{apex angle} turns out to be the most promising parameter of the Pyramid class. Two regimes appear. The first one corresponds to the classical case where the pupils images are completely separated, i.e. $2f\theta/D>1$. In this regime, all the performance criteria are stable regarding to $\theta$. Without modulation it corresponds to the classical \emph{flat} sensitivity and linearity range whatever the number of faces equals to. The second regime corresponds to the flattened Pyramids, i.e. $0<2f\theta/D<0.5$. On this range, the influence of $\theta$ is significant especially concerning the sensitivity and the linearity range. In particular, this parameter allows to choose where the gain in terms of sensitivity is maximal. Moreover, the small angle configurations allows to drastically improve the efficiency of the incoming photon without needing modulation. At the same time, one notes that few pixels are needed to do the wave front sensing. \noindent The next steps of this study consist in comparing all the WFSs of the Pyramid class on a unique \textbf{optical bench}. Following the results of Akondi et al.\cite{Akondi2013a} the technological device allowing to create all the Fourier masks will be a Spatial Light Modulator. Concerning the forthcoming numerical simulations, we plan to create an AO loop where all the optical parameters of the Pyramid WFS would be, as the deformable mirror, controlled and try to close it depending on the AO contexts. \noindent Finally, we mention that this study does not concern the impact of the high-orders spatial frequencies which are not corrected in a AO loop. However, it appears (Korkiakoski et al.\cite{Kork08,Korkk08}) that non-linearity caused by such residual phases has strong impacts on performance. Theoretical investigations will thus be led in this direction by extending our model in order to be able to handle both non null static reference phase and dynamic residuals phase.

1607.03269

1607

1607.06811_arXiv.txt

Chemical abundances are presented for 25 M31 globular clusters (GCs), based on moderately high resolution ($R=22,500$) $H$-band integrated light spectra from the Apache Point Observatory Galactic Evolution Experiment (APOGEE). Infrared spectra offer lines from new elements, of different strengths, and at higher excitation potentials compared to the optical. Integrated abundances of C, N, and O are derived from CO, CN, and OH molecular features, while Fe, Na, Mg, Al, Si, K, Ca, and Ti abundances are derived from atomic features. These abundances are compared to previous results from the optical, demonstrating the validity and value of infrared integrated light analyses. The CNO abundances are consistent with typical tip of the red giant branch stellar abundances, but are systematically offset from optical, Lick index abundances. With a few exceptions, the other abundances agree between the optical and the infrared within the $1\sigma$ uncertainties. The first integrated K abundances are also presented, and demonstrate that K tracks the $\alpha$-elements. The combination of infrared and optical abundances allows better determinations of GC properties, and enables probes of the multiple populations in extragalactic GCs. In particular, the integrated effects of the Na/O anticorrelation can be directly examined for the first time.

\label{sec:Intro} Integrated Light (IL) spectroscopy of globular clusters (GCs) provides valuable clues about the assembly histories of distant galaxies and their GC systems. An IL spectrum comes from an entire stellar population---integrated chemical abundances therefore represent flux-weighted averages from the individual stars observed in the IL spectrum. Despite the potential difficulties in modeling the underlying stellar populations, there are certain elements, spectral features, and/or wavelength regions that provide robust IL abundances (see, e.g., \citealt{SchiavonHB,Sakari2013}). Low and medium resolution ($R \lesssim 5000$) IL spectroscopy provides ages, metallicities, and abundances of the elements with the strongest spectral features (e.g., C, N, Mg; \citealt{Caldwell2011,Schiavon2013}), while higher resolution spectroscopy provides higher precision abundances of a wider variety of elements (including neutron capture elements such as Ba and Eu in the optical; \citealt{McWB,Colucci2009,Colucci2012,Sakari2015}). IL spectral observations have identified, among other things, possible metallicity bimodalities (\citealt{Perrett2002}, though \citealt{Caldwell2011} find no bimodality) and metallicity gradients (e.g., \citealt{Caldwell2011}) in M31's GC population, chemically-peculiar GCs in M31's outer halo that may have been accreted \citep{Sakari2015}, $\alpha$-deficiencies in distant GCs that are associated with dwarf galaxies \citep{Puzia2008}, and enhanced [$\alpha$/Fe] ratios in metal-rich GCs associated with the early type galaxy NGC~5128 \citep{Colucci2013}. IL spectroscopy has also provided insight into the nature of GCs themselves, through comparisons with Milky Way GCs \citep{Schiavon2012}, detections of anomalous abundances indicative of multiple populations (e.g. \citealt{Colucci2009,Colucci2014}, \citealt{Sakari2015}) and abundance correlations with cluster mass \citep{Schiavon2013}. Though previous IL observations have typically been at optical wavelengths ($\sim 3000-9000$ \AA), high resolution, infrared (IR) IL spectroscopy is now possible thanks to recent advances in infrared (IR) spectroscopy. In particular, the Apache Point Observatory Galactic Evolution Experiment (APOGEE) provides multi-object, high-resolution ($R=22,500$) spectroscopy with coverage in the $H$-band (from $1.51$ to $1.69\;\mu$m); this wavelength coverage has some significant advantages for IL spectroscopy. \begin{enumerate} \item {\it Insensitivity to hot stars.} IL spectra are composed of light from all the stars in a stellar population, including dwarfs and giants, hot and cool stars, etc. At blue wavelengths, contributions from hot horizontal branch (HB) stars complicate analyses (e.g. \citealt{SchiavonHB,Sakari2014}). Similarly, optical spectra are more sensitive to turnoff stars, and therefore require estimates of GC age. IR spectra are likely to be sensitive only to the brightest red giant branch (RGB) and asymptotic giant branch (AGB) stars, simplifying IL analyses. \item {\it Additional spectral lines.} The $H$-band offers different spectral lines than the optical. In particular, strong molecular lines of CN, CO, and OH enable determinations of C, N, and O abundances \citep{Smith2013}. There are CN indices in the optical, though they are in the blue and may be too weak in metal-poor clusters \citep{Schiavon2013}. The $H$-band also offers complementary lines to the optical, including additional Mg, Al, Si, Ca, and Ti lines. Stronger Al and Si lines can also be utilized in the IR than in the optical. \item {\it Opportunities to probe multiple populations in GCs.} The well-established chemical variations in Milky Way (MW) GCs (e.g., in Na/O and Mg/Al; \citealt{Carretta2009}) have been inferred to exist in extragalactic GCs because of their IL abundance ratios, notably high [Na/Fe] \citep{Colucci2014,Sakari2015}. The $H$-band offers detectable lines from elements that should vary within (at least some) GCs, including C, N, O, Mg, and Al. The ability to detect [O/Fe] and directly probe the Na/O anticorrelation makes the IR particularly valuable for extragalactic GC studies. \end{enumerate} \noindent However, IR IL spectroscopy also suffers from some disadvantages, compared to the optical. \begin{enumerate} \item {\it Line blending.} In metal-rich clusters molecular features dominate the $H$-band; as a result, abundances derived from lines in the $H$-band are sensitive to the abundances of C, N, and O. This blending is especially significant in spectra whose lines are already blended as a result of the cluster velocity dispersion. \item {\it Weak lines at low metallicity.} Many of the strong features in IL spectra become weaker in the more metal poor GCs, and may disappear entirely. For the most metal-poor clusters, $H$-band IL spectroscopy may therefore not provide abundances for as many elements as the more metal-rich GCs. \item {\it Sensitivity to evolved stars.} As stated above, the $H$-band is most sensitive to the brightest cluster RGB and AGB stars. The abundances are therefore sensitive to how the evolved AGB stars are modeled (in terms of the isochrones, the model atmospheres, the relative numbers of AGB stars, and stochastic sampling). \item {\it Lack of iron lines.} The high-resolution IL analyses of unresolved GCs that were developed by \citet{McWB}, \citet{Colucci2009,Colucci2011a,Colucci2014}, and \citet{Sakari2013,Sakari2015} rely on \ion{Fe}{1} lines to determine the parameters of the underlying stellar population (specifically the age and metallicity of an appropriate isochrone). However, there are few sufficiently strong Fe lines in the $H$-band (especially at low metallicities), and any detectable Fe lines may be blended with other features. \end{enumerate} \noindent Though IR IL spectroscopy may not be as informative as optical IL spectroscopy on its own, it offers valuable complementary information to the optical, {\it provided that IL analysis techniques are viable in the $H$-band.} For highly reddened objects the IR might also provide the only viable spectra for abundance analyses. This paper presents the first IR, IL detailed abundance analyses of GCs, utilizing targets in M31. The IR abundances are compared to those derived from optical lines (similar to the individual stellar analysis of \citealt{Smith2013}). The targets cover a wide range in metallicity and have all had previous high-resolution, detailed IL abundances analyses in the optical. These $H$-band IL spectra were observed during an ancillary program of the APOGEE survey, as described in Section \ref{sec:Observations}. The analysis techniques are introduced in Section \ref{sec:Analysis}; $H$-band abundances for 25 clusters are then presented (Section \ref{sec:Abunds}), along with new optical abundances for five GCs (Appendix \ref{appendix:OpticalAbunds}). The comparisons are discussed in detail in Section \ref{sec:Discussion}---particular emphasis is placed on the feasibility of future high-resolution IR IL analyses.

\label{sec:Conclusion} This paper has presented an $H$-band integrated light spectroscopic analysis of 25 bright GCs that are associated with M31. The target GCs span a wide range in metallicity (from $[\rm{Fe/H}] = -1.8$ to $-0.2$), a moderate range in total mass (from $\log \rm{mass} \sim 5.4$ to $\sim 6.5$), and a small range in age (from $\sim 6.5$ to $14$ Gyr). All the GCs were previously targeted for high-resolution optical spectroscopy, enabling a comparison between $H$-band and optical abundances. The primary results from this study are as follows. \begin{enumerate} \item The $H$-band offers a wide variety of spectral lines that complement the optical. In addition to the handful of additional \ion{Fe}{1} lines, the $H$-band offers intermediate and strong \ion{Mg}{1}, \ion{Al}{1}, and \ion{Si}{1} lines and weaker Na, Ca, Ti, and K lines. Molecular CO, CN, and OH features allow determinations of C, N, and O abundances. $^{12}$C/$^{13}$C ratios cannot be well constrained in these spectra, but could be measurable in higher quality spectra from metal-rich targets. \item The \ion{Fe}{1} lines in the $H$-band provide Fe abundances that are in excellent agreement with the optical abundances. Although there is a small offset ($\sim 0.05$) between the optical and $H$-band [\ion{Fe}{1}/H] ratios, this may reflect different NLTE corrections between the optical and IR (e.g., \citealt{GarciaHernandez2015}). Despite the small offset, the $H$-band lines agree with the optical trends in Fe abundance with line wavelength, reduced equivalent width, and excitation potential. However, the parameter ranges are smaller amongst the $H$-band lines, and it may not be possible to constrain the GC age without complementary optical data (photometric or spectroscopic). However, the $H$-band abundances ratios are relatively insensitive to GC age, at least for GCs older than $\sim 3$ Gyr. \item The $H$-band [C/Fe] and [N/Fe] abundances reflect typical tip of the RGB stellar abundances. However, they are systematically offset from the C and N abundances derived from the optical 4668 \AA \hspace{0.025in} Lick index. \item With a few exceptions, the abundances of Na, Mg, Al, Si, and Ca are in excellent agreement between the optical and the $H$-band. The $H$-band \ion{Ti}{1} abundances agree well with the optical \ion{Ti}{2} abundances, suggesting that the $H$-band lines may be less sensitive to NLTE effects than the optical ones. \item The $H$-band offers new [O/Fe] and [K/Fe] abundances that are not available in the optical. The K abundances trace Ca, as predicted from stars in the MW. \item With the detailed $H$-band abundances, multiple populations in extragalactic GCs can be explored in new detail. The integrated [Na/O] ratio is found to roughly correlate with cluster mass, suggesting that the relative numbers of ``second generation'' stars may increase with cluster mass. No convincing similar trend was found for [Mg/Al]. \item As expected for GCs associated with a massive spiral galaxy, the $H$-band [$\alpha$/Fe] ratios track MW field stars, demonstrating that $H$-band IL spectroscopy can be utilized for chemical tagging analyses of unresolved targets. \end{enumerate} Thus, $H$-band integrated light spectroscopy will be a valuable tool for studying more distant, unresolved stellar populations, particularly those that are highly reddened.

1607.06811

1607

1607.02108.txt

Coronal jets represent important manifestations of ubiquitous solar transients, which may be the source of significant mass and energy input to the upper solar atmosphere and the solar wind. While the energy involved in a jet-like event is smaller than that of ``nominal" solar flares and coronal mass ejections (CMEs), jets share many common properties with these phenomena, in particular, the explosive magnetically driven dynamics. Studies of jets could, therefore, provide critical insight for understanding the larger, more complex drivers of the solar activity. On the other side of the size-spectrum, the study of jets could also supply important clues on the physics of transients close or at the limit of the current spatial resolution such as spicules. Furthermore, jet phenomena may hint to basic process for heating the corona and accelerating the solar wind; consequently their study gives us the opportunity to attack a broad range of solar-heliospheric problems.

The wide variety of transient phenomena in the solar corona first became apparent in the 1970s with the discovery of coronal transients in ground-based, green-line observations \citep{1973SoPh...31..449D}; discovery of macro-spicules in Skylab EUV observations \citep{1975ApJ...197L.133B,1976ApJ...203..528W}; and the discovery of explosive events \citep{1980HiA.....5..557B}. These discoveries led to speculations on the role these transients, particularly coronal jets, play in the coronal heating and SW acceleration \citep{1978BAAS...10R.416B,1983ApJ...272..329B}. Coronal jets were seen by the U.S. Naval Research Laboratory (NRL)/UV telescope onboard the space shuttle in the 1980s and later by the Japanese spacecraft {\it{Yohkoh}} in the early 1990s. {\it{Yohkoh}}/SXT observations unveiled the largest, most energetic category of coronal jets \citep[e.g.,][]{1992PASJ...44L.173S,1992PASJ...44L.161S,1996PASJ...48..123S,1998SoPh..178..379S,2001ApJ...550.1051S}. Since then jet-like phenomena have occupied a center stage in coronal observational, theoretical, and state-of-the-art numerical analyses. Coronal jets are a near-ubiquitous solar phenomenon regardless of the solar cycle phase. They are particularly prominent in CHs (e.g., open magnetic field regions) because of the darker background. X-ray and EUV observations reveal their collimated, beam-like structure, which are typically rooted in CBPs. Their signature can be traced out to several Mm in X-ray/EUV observations, up to several solar radii in WL images \citep[e.g.,][]{1998ApJ...508..899W}, and also at $>1$~AU in in-situ measurements \citep[e.g.,][]{2006ApJ...639..495W,2006ApJ...650..438N,2008ApJ...675L.125N,2012ApJ...750...50N}. The unceasing improvements in spatial and temporal resolution of data recorded over the last three decades by different space missions (e.g., {\it{Yohkoh}}, {\it{SOHO}}, {\it{STEREO}}, {\it{Hinode}}, {\it{SDO}}, {\it{IRIS}}) provide unprecedented details on the initiation and evolution of coronal jets. The recent imaging and spectroscopic observations unveiled jet characteristics that could not be observed with lower spatio-temporal resolution (e.g., morphology, dynamics, and their connection to other coronal structures). Despite the major advances made on both observational and theoretical fronts, the underlying physical mechanisms, which trigger these events, drive them, and influence their evolution are not completely understood. Recent space missions (e.g., {\it{STEREO}}, {\it{Hinode}}, and {\it{SDO}}) represent important milestones in our understanding of the fine coronal structures, particularly coronal jets. The observations show that jets can be topologically complex and may contribute to the heating of the solar corona and the acceleration of the SW. The present review is the result of work performed by the ISSI International Team on ``Solar Coronal Jets". We, the authors, met at ISSI twice (March 2013 and March 2014) and had intense discussions on the nature of coronal jets, their triggers, evolution, and contribution to the heating and acceleration of the coronal and SW plasma, from both observational and theoretical point of views. We do not claim that this review is in any way exhaustive but it presents a thorough overview on the wealth of observations available from different space missions as well as state-of-the-art models of these coronal structures. The work we accomplished addressed many questions regarding coronal jets, but also left many others unanswered and raised several other outstanding issues for these prominent structures. Future missions with better observational capabilities along with the maturing of existing numerical codes will help address these questions and may lead to a yet better understanding of coronal jets and their role as a component of the magnetic activity of the Sun. In the present review, we mainly deal with observations from the {\it{SOHO}} era to the present. {\it{Yohkoh}}/SXT observations led to important insights and laid the seeds of major progress made during the later decades \citep[see, e.g.,][]{1992PASJ...44L.173S}. Chromospheric jets such as spicules may belong to the small-size end of jet phenomena and may be related to our topic. We feel, however, that such studies are beyond the objective of our review of coronal jets and should be excluded here. The vast literature on spicules and other chromospheric jets includes reviews of \citet{1968SoPh....3..367B,1972ARA&A..10...73B}, \citet{1974soch.book.....B}, \citet{1974IAUS...56....3M}, \citet{2000SoPh..196...79S}, and \citet{2012SSRv..169..181T}.

Imaging and spectroscopic observations over the last two decades have provided unprecedented insights into the formation and evolution of solar coronal transients, particularly coronal jets. Recent space missions, such as {\it{Hinode}}, {\it{STEREO}}, {\it{SDO}}, {\it{IRIS}} and are instrumental in advancing our understanding of this phenomenon. Instrument improvement in terms of both spatial and temporal resolution and also temperature coverage are key in the numerous discoveries made concerning the different facets of coronal jets. Thanks to the multiple discoveries made using high quality observations (both remote sensing and in situ), it is now widely believed that coronal jets play an important role in the multi-scale solar activity, coronal heating, and the contribution to the SW. The Feature Finding Team (FFT) of {\it{SDO}} developed dedicated computer tools that allow identifying and characterizing coronal features including coronal jets. This has the potential to build massive jet statistics all over the solar disk. This would hopefully give us truly global statistics of the energy and mass contents of jets and their role in the energy and mass supply of the solar atmosphere However, there are still many aspects of coronal jets that remain ambiguous and need further investigation. It is still unclear whether the scale size spectrum of coronal jets extends to scales much smaller than the spatial resolution available now. Do they extend to the nano-flare scales where it is believed that the contribution to the coronal heating could be significant? The contribution of jets to the SW and to the population of energetic particles is still unclear as well as their role in driving the formation and evolution of other coronal structures such as plumes and chromospheric features such as spicules. An area of improvement in the study of coronal jets that need to be deepened is spectroscopy, which provides insights into the plasma properties of jets. This aspect is still in its infancy compared to imaging. The latter needs further improvements in terms of spatial resolution. The extensive jet modeling in the past decade or so has achieved many successes including the detailed representation of the jet morphology and dynamics that match observations of jets well. In addition, we have achieved the successful modeling of the transition between standard and blow-out jets as well as the ability to create recurrent jets and jet/plume structures. Jets have been produced both in the open field of CHs and the large-scale closed field of solar ARs. The scenario for producing jets that has been most widely explored in the simulations is flux emergence \citep{2008ApJ...673L.211M, 2009ApJ...704..485T, 2013ApJ...771...20M,2013ApJ...769L..21A}, but increasing attention has been given to the instability-onset scenario \citep{2009ApJ...691...61P,2010ApJ...714.1762P,2015A&A...573A.130P}. The mechanism of flux cancellation has not been explored with MHD simulations, while a small study of flux cancellation has been performed in a magnetofrictional simulation. Recent observations \citep[e.g.,][]{2014SoPh..289.3313Y} demonstrate that the flux cancellation mechanism may be as important as the emergence. Most current simulations lack a full thermodynamic treatment and do not include thermal conduction or radiation effects. These ingredients are clearly important for reproducing the observed plasma properties and understanding the emission and spectral observations of jets. As a consequence the MHD simulations so far either under- or overestimate the temperature of jets. As more plasma diagnostics of coronal jets become available through analysis of {\it{Hinode}}/EIS and {\it{IRIS}} data, we need to create increasingly realistic MHD simulations of jets. Current MHD simulations only deal with idealized boundary and initial conditions. Hence, the ultimate goal is to develop data-constrained, and eventually, data-driven MHD simulations with useful energy equations to model observed events. Other investigations, both observational and theoretical, should clarify whether mini-filament eruptions play a larger role than previously recognized in jet and jet-bright-point formation \citep{sterling.et15}. We believe that future missions such as NASA's Solar Probe Plus and ESA's Solar Orbiter will provide further insights into the physics of coronal jets and the related phenomena. For instance, Solar Probe Plus will fly through coronal structures including jets, which would provide close by imaging observations as well as in situ measurements of these features. Solar Orbiter's above the ecliptic observations will provide unprecedented view of the solar poles where coronal jets are prominent. This includes magnetic field measurements, spectroscopy, imaging, and in situ measurements.

1607.02108

1607

1607.00380_arXiv.txt

High-energy cosmic rays can be accelerated in clusters of galaxies, by mega-parsec scale shocks induced by accretion of gas during the formation of large-scale structure, or by powerful sources harbored in clusters. Once accelerated, the highest energy particles leave the cluster via almost rectilinear trajectories, while lower energy ones can be confined by the cluster magnetic field up to cosmological time and interact with the intracluster gas. Using a realistic model of the baryon distribution and the turbulent magnetic field in clusters, we studied the propagation and hadronic interaction of high-energy protons in the intracluster medium. We report the cumulative cosmic ray and neutrino spectra generated by galaxy clusters including embedded sources, and demonstrate that clusters can contribute a significant fraction of the observed IceCube neutrinos above 30 TeV while remaining undetected in high-energy cosmic rays and $\gamma$ rays for reasonable choices of parameters and source scenarios.

\label{sec:introduction} The IceCube Observatory has observed the first high-energy astrophysical neutrinos \citep{Aartsen:2013jdh, Aartsen:2013bka, Aartsen:2013uuv}. The all-flavor diffuse neutrino flux is reported to be $\Phi_\nu=2.06\times10^{-18}\,(E_\nu/10^5\,\rm GeV)^{-2.46}\,\rm GeV^{-1}cm^{-2}sr^{-1}s^{-1}$ for the energy range $25\,\rm TeV < E_\nu < 1.4\,\rm PeV$ \citep{2015PhRvD..91b2001A}. Searches for small-scale anisotropies from the three-year IceCube data do not see any significant clustering or correlations \citep{Aartsen:2014ivk}. The origin of these high-energy neutrinos remains unclear \citep{MuraseReview}. Accretion of gas during the formation of the large-scale structure can give rise to mega-parsec scale shocks that accelerate high-energy cosmic rays \citep{2000ApJ...542..608M, Ryu:2003cd}, and even ultra-high energy cosmic rays (UHECRs) if the medium around the shocks contain heavy nuclei \citep{Inoue07}. Numerical simulations \citep{Vazza:2014usa, 2014ApJ...785..133H, 2015ApJ...800...60M} also suggest the possibility of stochastic particle acceleration in large-scale structures, though the flux level of cosmic rays depend on the Mach number and the location of shocks \citep{Vazza:2014usa}. Powerful sources harbored in the galaxy clusters can also accelerate particles to high energies. Many plausible candidate sources have been proposed in literature, including steady sources like active galactic nuclei (AGN) (e.g. \cite{PhysRevLett.66.2697, Winter:2013cla, Murase:2014foa, Dermer:2014vaa}), and transients like gamma-ray bursts (GRBs) (see \cite{0034-4885-69-8-R01} for review), fast-spinning newborn pulsars \citep{Blasi00, Fang:2012rx, Fang:2013vla}, magnetars \citep{Arons:2002yj, 2009PhRvD..79j3001M}, and blazar flares \citep{Farrar:2008ex}. Clusters are famously known as cosmic ray reservoirs, due to their ability to confine cosmic rays with turbulent magnetic fields up to cosmological time (\citet{1996SSRv...75..279V, 0004-637X-487-2-529}, also see \cite{2015ASSL..407..557B} for a recent review). Hence the accelerated cosmic rays have a good chance to interaction with the intracluster medium (ICM), leading to the production of secondary neutrinos and $\gamma$ rays. A cosmic ray reservoir scenario is favored to explain the absence of detection of neutrinos above a few PeV, especially around the Glashow resonance at 6.3 PeV. This is because rather than otherwise a mysterious stop in the cosmic ray spectrum, a ``cosmic ray reservoir" scenario naturally introduces a spectral softening caused by the faster escape of the higher-energy cosmic rays in a magnetized source environment. Many analytical and semi-analytical works were conducted to calculate these secondary fluxes \citep{1995PhRvL..75.3052D, 1996SSRv...75..279V, 0004-637X-487-2-529, 2006PhRvD..73d3004D, Murase:2008yt, 2008ApJ...687..193W, Murase:2013rfa}. However most of the analytical approaches adopted an overly simplified modeling of the ICM by assuming both uniform gas distribution and uniform magnetic field. \cite{1998APh.....9..227C} took into account the density profile and mass function of galaxy clusters, yet still assumed a uniform magnetic field within each cluster and the same cosmic ray diffusion length for all groups of clusters with different masses. Numerical propagation of cosmic rays in more realistic three-dimensional cluster magnetic fields were explored by \cite{2004APh....22..167R} and \cite{Kotera09}, but the studies were limited to neutrinos from a single type of clusters. Numerical modeling of the non-uniform gas distribution and magnetic field structure is crucial in an environment like the ICM, because both duration and location of the confinement of the charged particles impact the interaction rate, and thus an accurate computation of the neutrino production. Moreover, although massive clusters could be brighter cosmic ray and neutrino sources, they are far less common than medium-sized clusters. It is thus unclear how the mass and redshift dependence of cluster number density would impact the total neutrino spectrum. It is also unknown which group of clusters would dominate PeV cosmic rays and thus be more relevant for the detected neutrinos. The latter point is also important for revealing the possibility of pinpointing sources with the increasing statistics of IceCube as well as future experiments \citep{Ahlers:2014ioa, Fang:2016hyv}. \cite{Bechtol:2015uqb} pointed out that new studies of the blazar flux at gamma rays above 50 GeV result in a lower residual non-blazar component of the extragalactic gamma-ray background. This puts a tight constraint on neutrino sources that are transparent to gamma rays \citep{Murase:2015xka}, especially those which produce neutrinos through hadronuclear (pp) interactions (as in galaxy clusters \citep{Kotera09}). However, this constraint is drawn based on the assumption that the pp scenario has a $E^{-2}$ spectrum extended below $\sim 10$ TeV, which is not necessarily valid for all source types. For example, particles accelerated by fast-spinning newborn pulsars \citep{Blasi00, Fang:2012rx, Fang:2013vla} can have a spectrum index less than 2, and cosmic rays accelerated in the AGN \citep{2012ApJ...749...63M, 2012ApJ...755..147D} can present a cutoff at low energies due to the confinement of the source environment. In this paper, we investigate high-energy cosmic rays and neutrinos from clusters, by numerically propagating cosmic rays down to TeV in galaxy clusters with a wide range of masses and redshifts. We limit the uncertainties of our results with a more realistic modeling of the ICM gas and the turbulent cluster magnetic field than some of the previous works, and by adopting the mass accretion rates, cluster baryon fraction, and halo mass function as constrained by cosmological observations. We report the integrated cosmic ray and neutrino spectra from the entire cluster population in two scenarios: \begin{enumerate} \item the {\it accretion shock scenario}: cosmic rays are accelerated by the cluster accretion shocks and injected at the outskirts of clusters; \item the {\it central source scenario}: cosmic rays are accelerated by sources at the centers of the clusters. \end{enumerate} Our results demonstrate that neutrinos from cluster accretion shocks could contribute $\lsim 20\%$ of the IceCube flux; however, if bright astrophysical sources inside the galaxy clusters can accelerate particles with an injection spectrum index below 2, clusters could reproduce both the spectrum and flux of IceCube neutrinos above $30$ TeV, while remaining consistent with the measurements of high-energy cosmic rays and $\gamma$ rays.

\label{sec:discussion} Unlike other astrophysical sources, galaxy clusters offer a unique environment for TeV-PeV neutrino production through efficient acceleration and confinement of high-energy cosmic rays. By propagating particles in three-dimensional turbulent magnetic fields and recording their interactions with the ICM gas, we find that the integrated neutrino flux from the cluster accretion shocks could account for $\le 20\%$ of the IceCube detections, while neutrinos produced by the interaction of cosmic rays from powerful sources hosted by clusters could explain both the flux and spectrum of the IceCube data above 30 TeV, if the injection spectrum is harder than 2. The lack of neutrino clustering around known sources also fits well with the cluster production model. In addition, the high-energy cosmic rays that succeed in escaping from the clusters contribute less than 20\% of the observed cosmic ray flux around and below the ankle, and may explain the second knee feature when galactic and ultrahigh energy cosmic ray accelerators contributions are added to the clusters' contribution. Our results demonstrate that when taking into account the effect of cosmic ray confinement and the mass dependence of the cluster number density, the integrated neutrino spectrum would conserve the injection spectrum around TeV - 100 TeV, but become much steeper above PeV. If high-energy sources harbored in the galaxy clusters could produce cosmic rays with spectrum harder than $E^{-2}$, the tension between the neutrino-associated GeV $\gamma$ rays and the Fermi measurement of isotropic diffusive $\gamma$-ray background \citep{0004-637X-799-1-86,Murase:2015xka} can be alliviated. Also notice that in the cluster scenario, the dominant contribution comes from sources beyond $z\sim0.3$ (as suggested by Fig.~\ref{fig:PhiNeu}). Thus, $\gamma$ rays with $E\gsim$0.1 TeV are attenuated by the time they arrive on Earth, through interactions with photons of the extragalactic background light (EBL) and the CMB that have a non-trivial optical depth, $\tau_{\gamma\gamma}(z\sim0.3)> 1$ \citep{2006ApJ...648..774S}. The search for the first $\gamma$ rays from galaxy clusters is still ongoing. No significant spatially extended $\gamma$-ray emission from the nearby galaxy clusters was found in four years of Fermi-LAT data, establishing limits on the cosmic ray to thermal pressure ratio, $X_{\rm CR}$, to be below 1.4\% \citep{2014ApJ...787...18A}. The parameter $f_{\rm CR}$ of our model can be translated to $X_{\rm CR} $ by $X_{\rm CR} = P_{\rm CR} / P_{\rm th}\sim 0.6\%\,f_{\rm CR,-2}\, (\dot{M}\,t_{\rm cr, conf}/M)$, where $P_{\rm CR}\approx (1/3)\,f_{\rm CR}(G\dot{M}\rho_{\rm ICM}\,t_{\rm cr,conf}/r_{\rm vir})$ is the cosmic ray pressure after an accumulation of particles for $t_{\rm cr, conf}$, and $P_{\rm th } = n_{\rm ICM} (GM\mu m_p/ 2\,r_{\rm vir})$ is the thermal pressure. The $f_{\rm CR} <2\%$ suggested by our model is thus consistent with the constraint from $\gamma$-ray searches. However, the Fermi limit would constrain cluster scenarios with $f_{\rm CR}>2\%$, corresponding to an injection index $\alpha>2.1$. In the central source scenario (bottom panel of Fig.~\ref{fig:PhiNeu}) we showed a benchmark case with an injection spectrum index $\alpha = 1.5$ and maximum cosmic ray energy $E_{\rm max} = 50\,\rm PeV$. The results are solid under moderate adjustments on either $\alpha$ or $E_{\rm max}$ (including adding a mass dependence). However, $\alpha$ cannot be as large as $\sim$ 2, otherwise the $\gamma$-ray counterpart will clearly exceed the Fermi measurements of the isotropic $\gamma$-ray background. Furthermore, if $E_{\rm max}\ll 50 \,\rm PeV$, the injected cosmic rays wouldn't be energetic enough to produce the PeV neutrinos. On the other hand, if $E_{\rm max}\gg 50$ PeV in all clusters, the resulted neutrino spectrum could overshoot the null bins at 5 - 10 PeV. In our simulations we assumed a central magnetic field strength of $3\,\mu$G. In case of a smaller central magnetic field strength $B_0$, highest energy cosmic rays would have a higher chance to leave the system due to the weaker magnetic field and less interaction materials in the environment. Lower energy cosmic rays would be less impacted, since they are confined to a relatively small volume. Hence the overall neutrino spectrum would be expected to have a lower flux but a softer spectrum. Conversely, stronger $B_0$ would lead to a higher neutrino flux due to the intenser confinement. Neither of our scenarios could account for the IceCube data below 30 TeV. Although majority of the neutrinos are expected to come from extragalactic sources, Galactic sources could potentially contribute some fraction of the flux \citep{Ahlers:2015moa}, especially below $\sim 200 \,\rm TeV$ \citep{MuraseReview}. It is also possible that other types of sources could contribute to this energy range. Our conclusion on the accretion shock scenario is consistent with that from \cite{2015A&A...578A..32Z}, though the two works have distinctive approaches. While \cite{2015A&A...578A..32Z} derived the neutrino luminosity from a scaling between radio and gamma-ray luminosities, we directly simulated the particle propagation to obtain the neutrino spectrum. \cite{2015A&A...578A..32Z} did not assume any cosmic ray spectral steepening due to the escape of high-energy cosmic rays, whereas in our work this cutoff was shown as a natural result from the energy-dependent transportation, and crucial for the survival of the central source scenario. On the other hand, \cite{2014MNRAS.438..124Z, 2015A&A...578A..32Z} indicated that the neutrino contribution from clusters could be further limited due to the fact that not all clusters are expected to produce hadronic emissions. The same constraint could also apply to our results in the accretion shock scenario. Although consistent with current $\gamma$-ray and radio detection limits, the cluster scenario could be robustly tested by the growing statistics of IceCube as well as $\gamma$-ray searches of galaxy clusters in the near future. If the cluster contribution to the diffusive $\gamma$-ray background and the cosmic ray to thermal pressure ratio can be further constrained by the $\gamma$-ray observations, a significant contribution to IceCube neutrinos from galaxy clusters would be ruled out. On the other hand, as our model predicts a dominant neutrino flux from the most massive clusters above $\sim$5 PeV, future detection of strong anisotropy in this energy regime would provide a firm support to the cluster scenario.

1607.00380

1607

1607.07872_arXiv.txt

{ We study the effects of the non-attractor initial conditions for the canonical single-field inflation. The non-attractor stage can last only several $e$-folding numbers, and should be followed by hilltop inflation. This two-stage evolution leads to large scale suppression in the primordial power spectrum, which is favored by recent observations. Moreover we give a detailed calculation of primordial non-Guassianity due to the ``from non-attractor to slow-roll'' transition, and find step features in the local and equilateral shapes. We conclude that a plateau-like inflaton potential with an initial non-attractor phase yields interesting features in both power spectrum and bispectrum.}

\label{intro} Inflationary scenario is the most popular paradigm for the very early universe which successfully resolves the puzzles of standard big bang cosmology~\cite{Starobinsky:1980te,Fang:1980wi,Guth:1980zm,Sato:1980yn,Linde:1981mu, Albrecht:1982wi}. Conventionally, inflation is driven by the potential of a slowly rolling canonical scalar field, the inflaton, and this evolution is described by the smallness of the so-called slow-roll parameters. It predicts a nearly scale-invariant power spectrum of the primordial curvature perturbation, which has been verified to high precision by the latest measurements of cosmic microwave background (CMB) temperature anisotropies~\cite{Ade:2015xua,Ade:2015lrj}. From the observational aspects, however, there are several issues remaining to be settled. For example, the primordial non-Gaussianity, as a powerful tool to distinguish among various models of inflation and alternatives, has not been detected. Also the recent CMB observations seem to indicate a relative power suppression on low multipoles~\cite{Ade:2015lrj}, which deviates from the prediction of naive single-field slow-roll inflation. These unresolved anomalies may well indicate that, upon further supply of more accurate observational data, our fiducial model of canonical single-field slow-roll inflation is not sufficient, and that we indeed observe the trail of the earlier phase of inflation or its alternatives. One theoretical direction to accommodate the observed CMB anomalies is to abandon the {\em assumption} of slow-roll, attractor evolution of the inflaton from beginning to end. Indeed, a few models of non-attractor phase of inflation have been suggested and studied, with one representative model being the ultra-slow-roll inflation~\cite{Tsamis:2003px,Kinney:2005vj}. In this model, the first slow-roll parameter remains very small but the second one is $\mathcal{O}(1)$ during inflation, leading to the growth of the curvature perturbation after horizon exit rather than remaining constant (see also~\cite{Motohashi:2014ppa}). Then it is realized that the non-attractor evolution can lead to a large bispectrum in the squeezed configuration even in single-field inflation models with a Bunch-Davies vacuum~\cite{Namjoo:2012aa,Martin:2012pe,Chen:2013aj,Chen:2013eea}, which violates the non-Gaussianity consistency relation~\cite{Maldacena:2002vr,Creminelli:2004yq}. This behavior of the curvature perturbation is the same as the one in the matter bounce cosmology, where the matter dominated contraction can also be seen as a ``non-attractor'' stage~\cite{Cai:2009fn, Cai:2014bea, Brandenberger:2016vhg}. Moreover, the latest CMB observations give rise to an upper bound for the tensor-to-scalar ratio as $r<0.07$~\cite{Ade:2015lrj, Array:2015xqh}. This result favors the inflation models with a very flat, plateau-like potential, such as the $R^2$ inflation~\cite{Starobinsky:1980te} and the recent $\alpha$-attractors~\cite{Kallosh:2013yoa}. There have been more and more discussions about the plateau-like potentials constructed from fundamental theories and their conceptual issues, such as eternal inflation~\cite{Ijjas:2015hcc}. The slow-roll trajectories of these models usually begin with a nearly vanishing field momentum at large field values. However, this needs not necessarily be the case and an initial non-attractor stage with relatively large field velocity can provide preferred observational consequences. For example, a preceding fast-roll stage on a steeper fraction of the potential may well provide a large field velocity~\cite{Jain:2008dw}. Therefore it is interesting to study the effects of these more general initial conditions. In this work, we phenomenologically study the non-attractor beginning of inflation on the plateau-like potential. We first note that the duration of canonical non-attractor inflation is limited to several $e$-folding numbers, and a hilltop potential with slow-roll evolution is needed to render it complete. It is shown that because of the relaxed initial condition, the primordial spectrum is modulated to generate features which may explain the low-$\ell$ suppression of power in the CMB temperature anisotropies. Moreover we find that, the ``non-attractor to slow-roll'' evolution retains the large non-Gaussianity generated in the first stage but modified the amplitude. Meanwhile there are also step features in both the local and equilateral limits of the bispectrum. The article is organized as follows. In Section~\ref{sec:non-attractor}, we quickly review the non-attractor inflation and the calculation of the primordial curvature perturbation during this stage, which leads to the limited $e$-folding number. In Section~\ref{sec:slow-roll}, we first discuss which type of slow-roll models could be connected to the non-attractor phase, and then study the background evolution with a relaxation stage and the resulting power spectrum. In Section~\ref{sec:nG}, we use the formalism to calculate the bispectrum and see the interesting features caused by the non-attractor initial condition. In Section~\ref{sec:plateau}, we briefly remark the implications on the inflation models with plateau-like potentials. In Section~\ref{sec:discussions}, we give some discussion and conclude this work.

\label{sec:discussions} There are two interesting directions regarding the non-attractor phase during inflation. First, non-attractor inflation models have aroused a lot theoretical concerns, since in this case the curvature perturbation grows after horizon-exit and the non-Gaussianity consistency relation is violated. Second, from the observational aspects, the latest data favors a class of inflation models, in which the inflaton rolls down a very flat, plateau-like potential. Motivated by these, we study the observable effects of the non-attractor initial conditions of inflation. In this work, we first review the non-attractor inflation using a toy model with a canonical scalar field, and find that this model can last for only several $e$-folding numbers thus a second stage of inflation should follow. Then to be consistent with the recent observations, the following slow-roll phase of inflation should be driven by hilltop potentials. This model construction is phenomenologically similar to the plateau-like potentials. Next, the power spectrum is calculated and we find that the non-attractor beginning leads to large-scale suppression in the CMB $TT$ spectrum, favored by recent observations. We also work out the bispectrum, and the result shows that large non-Gaussianity generated in the non-attractor stage will not survive the following evolution, but the transition process from non-attractor to slow-roll leads to step features in both the squeezed and equilateral limits of the bispectrum. These features from the non-attractor beginning of inflation will be interesting for future CMB observations. Finally we remark that, this work only cares for inflation driven by a single canonical field, which is the simplest case for the non-attractor models. Thus for future researches, it will be interesting to further study the non-attractor initial conditions for more general models with k-essence field. Meanwhile we have assumed the background dynamics is driven by the potential energy all the time. However, it is also tempting to abandon this assumption and consider a kinetic-dominated fast-roll stage before inflation. We expect that further relaxing the initial condition of inflation may result in more interesting features. \subsection*{Acknowledgements} We are grateful to Razieh Emami, Chunshan Lin, Junyu Liu, Jerome Quintin, Misao Sasaki, Tao-Tao Qiu, Kai Wang and Yi Wang for helpful discussions. We also thank KITPC for hospitality during the workshop ``Early Universe, Cosmology and Fundamental Physics", where this work was initiated. YFC, DGW and ZW are supported in part by the Chinese National Youth Thousand Talents Program and by the USTC start-up funding (Grant No.~KY2030000049) and by the National Natural Science Foundation of China (Grant No.~11421303). JG acknowledges support from the Korea Ministry of Education, Science and Technology, Gyeongsangbuk-Do and Pohang City for Independent Junior Research Groups at the Asia Pacific Center for Theoretical Physics. JG is also supported in part by a Starting Grant through the Basic Science Research Program of the National Research Foundation of Korea (2013R1A1A1006701) and by a TJ Park Science Fellowship of POSCO TJ Park Foundation.

1607.07872

1607

1607.00725_arXiv.txt

{The $\gamma$-ray BL Lac object OJ\,287 is known to exhibit inner-parsec ``jet-wobbling'', high degrees of variability at all wavelengths and quasi-stationary features including an apparent ($\approx 100^{\circ}$) position angle change in projection on the sky plane.} {Sub-$50$ micro-arcsecond resolution 86\,GHz observations with the global mm-VLBI array (GMVA) supplement ongoing multi-frequency VLBI blazar monitoring at lower frequencies. Using these maps together with cm/mm total intensity and $\gamma$-ray observations from \emph{Fermi}/LAT from 2008-2014, we aimed to determine the location of $\gamma$-ray emission and to explain the inner-mas structural changes. } {Observations with the GMVA offer approximately double the angular resolution compared with 43\,GHz VLBA observations and allow us to observe above the synchrotron self-absorption peak frequency. Fermi-LAT $\gamma$-ray data were reduced and analysed. The jet was spectrally decomposed at multiple locations along the jet. From this we could derive estimates of the magnetic field using equipartition and synchrotron self-absorption arguments. How the field decreases down the jet allowed an estimate of the distance to the jet apex and an estimate of the magnetic field strength at the jet apex and in the broad line region. Combined with accurate kinematics we attempt to locate the site of $\gamma$-ray activity, radio flares and spectral changes.} {Strong $\gamma$-ray flares appeared to originate from either the ``core'' region, a downstream stationary feature, or both, with $\gamma$-ray activity significantly correlated with radio flaring in the downstream quasi-stationary feature. Magnetic field estimates were determined at multiple locations along the jet, with the magnetic field found to be $\geq$\,1.6\,G in the ``core'' and $\leq$\,0.4\,G in the downstream quasi-stationary feature. We therefore found upper limits on the location of the VLBI ``core'' as $\lesssim$\,6.0\,pc from the jet apex and determined an upper limit on the magnetic field near the jet base of the order of thousands of Gauss. } {}

Radio loud active galactic nuclei (AGN) feature highly energised relativistic jets which are likely produced by the conversion of gravitational energy around a central super-massive black hole (SMBH) \citep{BZ77,BP82}. Blazars are a subclass of AGN with the jet direction being nearly parallel to our line of sight \citep{Urry95}. This causes relativistic effects including apparent superluminal motion, reduction of variability timescales and the apparent quasi-periodic changes of inner jet orientation that we refer to as jet ``wobbling''. This ``wobbling'' is thought to to be caused either by geometric effects \citep[e.g.,][]{jor05,bach05} or due to binary black hole procession \citep[e.g.,][]{binary_ref}. The BL Lac object OJ\,287 \cite[$z$=0.306,][]{z_ref} is a well studied blazar, harbouring a SMBH with widely varying mass estimates of $~4 \times 10^{8}-~1.8 \times 10^{10}$\,M$_{\odot}$ and exhibiting quasi-periodic flaring that has been suggested as due to a binary black hole system \citep{mass_ref,liu2002,binary_ref_2}. \\ \begin{table*}[th] \caption{Overview of VLBI observations of OJ\,287} \begin{center} \begin{tabular}{ccccccc} \hline \hline Epoch & Frequency & Participating Stations & Beam & Position Angle & Recording rate & Polarisation \\ & [GHz] & & [maj:min mas] & [$^{\circ}$] & [Mbit/s] & \\ \hline 2008.77 & 86.2 & All & 0.211 ; 0.047 & -9.3 & 512 & Dual$^{3}$ \\ 2009.36 & 86.2 & All & 0.219 ; 0.051 & -2.5 & 512 & Dual$^{3}$ \\ 2009.77 & 86.2 & All & 0.221 ; 0.045 & -2.6 & 512 & Dual$^{3}$ \\ 2010.35 & 86.2 & All & 0.269 ; 0.056 & -2.7 & 512 & Dual$^{3}$ \\ 2011.36 & 86.2 & All & 0.245 ; 0.047 & -5.6 & 512 & Dual$^{3}$ \\ 2011.76 & 86.2 & All & 0.255 ; 0.078 & 3.2 & 512 & Dual$^{3}$ \\ 2012.38 & 86.2 & All$^{2}$ & 0.230 ; 0.063 & 22.0 & 512 & Dual$^{3,4}$ \\ 2007.45-2013.57 & 43.13 & VLBA & 0.351 ; 0.145$^{1}$ & -2.9 & 512 & Dual \\ 2008.70-2012.39 & 15.36 & VLBA & 0.891 ; 0.379$^{1}$ & -5.5 & 512 & Dual \\ \hline \multicolumn{6}{l}{\textsuperscript{}\footnotesize{$^{1}$ Beam sizes are indicative only and will depend on uv coverage. $^{2}$ Yebes participated. }} \\ \multicolumn{6}{l}{\textsuperscript{}\footnotesize{$^{3}$ Onsala only supported LCP. $^{4}$ Yebes only supported LCP.}} \end{tabular} \end{center} \label{epochs2} \end{table*} The jet kinematics, light curves and polarisation properties of OJ\,287 have been recently studied by \citet{agudo11,agudo12}, with the position angle (PA) of the jet axis appearing to change by $\approx$\,$100^\circ$ between 2004 and 2006. Gamma-ray emission was suggested to be correlated with mm-radio flaring and placed at least 14 pc away from the central engine, largely in agreement with spectral energy distribution (SED) modelling by \citet{kushwaha13}. Recent flaring activity in 2011-2012 was analysed by \citet{sawada15} using 22\,GHz VLBI maps. They interpreted the $\gamma$-ray flaring as possibly being a new jet component that is unresolved at 22\,GHz. \\ Very long baseline interferometry (VLBI) has long been used to provide high angular resolution images of blazars, with angular resolutions of $\approx$0.1--0.2\,milli-arcseconds (mas) achievable with 43 GHz VLBI. Visible structural variations can occur within weeks to months, requiring high cadence monitoring. Such monitoring programs include the Monitoring Of Jets in AGN with VLBA Experiments (MOJAVE) program at 15\, GHz \citep{MOJAVE_ref} and the VLBA-BU-BLAZAR 43 GHz blazar monitoring program \citep{mar08}. A subset of these sources has been observed approximately semi-yearly at 86\,GHz using the Global mm-VLBI Array (GMVA) from 2008 until now. It has been proposed by \citet{DalyMarscher88}, \citet{Marscher08Rev}, \citet{Cawthorn13} and others that the mm-wave ``core'' in VLBI images could be the first of a series of recollimation shocks produced when a jet becomes under-pressured compared to its surrounding medium. At lower frequencies, the ``core'' is usually assumed to be a $\tau=1$ surface where the radiation begins to be self absorbed. Global 3\,mm VLBI using the GMVA allows the imaging of regions above the self-absorption turnover frequency with angular resolution approaching $\approx$40\,$\mu$as. \\ Since the launch of EGRET, on board the Compton Gamma Ray Observatory \citep{egret}, $\gamma$-rays have been known to be produced in AGN, but the site of their production remains elusive \citep{fichtel94,jor01}. With the launch in 2008 of the \emph{Fermi Gamma-ray Space Telescope} and the Large Area Telescope instrument on board (\emph{Fermi}), long-term $\gamma$-ray light curves of many ($>$\,$10^{3}$) AGN have been observed \citep{abdo10,2fgl}. Sites for $\gamma$-ray production are proposed to be either within the Broad Line Region (BLR) close to the central engine \citep{blandford95} or further along the jet in recollimation shocks or other jet features \citep{mar14}. The observational evidence is conflicting with some favouring the former scenario \citep[e.g.,][]{tav10,rani13b} and others the latter \citep[e.g.,][]{jor10,rani13a,fuhrmann14}. While TeV emission is not detected from all blazars, it is interesting to note that the comparable source with a similar redshift 0716+714, exhibits TeV emission, while OJ\,287 does not \citep{tash08,anderhub09,wang13}.\\ Here, we aimed to further test this scenario, using recent 15 and 43 GHz data and the semi-annual GMVA observations at 86\,GHz to derive kinematic properties and estimate magnetic field strengths in individual VLBI components. In Section 2, we present the data obtained and the methods to reduce and analyse the data. In Section 3, we present our observational results and Section 4 contains a deeper analysis using a spectral decomposition to compute magnetic field strengths at multiple locations in the jet. We then devise a method to estimate the location of jet features relative to the jet apex. In Section 5, we present the interpretation of these results and discuss them in the context of prevailing theories. In Section 6, we present our conclusions and outlook for the future. Dates throughout the paper are presented in decimal years. A linear scale of 4.64 pc/mas and a luminosity distance $D_{L}$ of 1.63 Gpc, at the source redshift of z=0.306 was adopted with standard cosmological parameters of $\Omega_{m} = 0.302$, $\Omega_{\lambda} = 0.698$ and $H_{0} = 68$ km s$^{-1}$ Mpc$^{-1}$ \citep{cosmo,spe07}. \\

We have used multi-frequency VLBI data at 2\,cm, 7\,mm and 3\,mm, total intensity radio data at 2\,cm, 7\,mm, 3\,mm, 1\,mm, 0.85\,mm and $\gamma$-ray data from the \emph{Fermi}/LAT space telescope to perform a detailed kinematic and spectral analysis of the highly variable BL Lac OJ\,287. We have used the 2\,cm MOJAVE, 7\,mm VLBA-BU-BLAZAR and 3\,mm GMVA data to determine the spectrum and hence estimates of the magnetic field at multiple locations down the jet. We combine this with kinematics derived from 3\,mm and 7\,mm data to determine the location of radio and $\gamma$-ray flaring events within the jet. We have found: \begin{enumerate} \item OJ\,287 exhibits two stationary features, components C (``core'') and S (``stationary feature''), that have very similar properties, with the ``core'' exhibiting an optically thin spectrum on two occasions. Both are interpreted as standing shocks. We postulate that the stationary feature moves on the sky relative to other components and we postulate that the ``core'' could also. \item The $\approx$\,100$^{\circ}$ PA change reported by \citet{agudo11} resulted in a radically different ejection position angles and trajectories. Recent data suggests that these values may be returning to their pre-2006 values. We suggest an alternate interpretation of this behaviour as due to a large disturbance at the jet base changing the jet pressure causing the location of downstream standing shocks to shift their locations in the jet and changing the viewing angle and hence the Doppler factor. \item Radio flaring activity was found in both the ``core'' and stationary feature. A large mm-wave radio flare (R2) was found to be a superposition of flares in both the ``core'' and stationary feature. Cross correlation analysis of $\gamma$-ray flaring found that flaring likely correlated with radio flaring in both the ``core'' and stationary feature as flaring coincides with the passage of components through both features. \item The magnetic field, as derived from SSA and from equipartition, decreases by $\sim80$ between the ``core'' and stationary feature. This allowed us to derive an estimate of the distance between the mm-wave ``core'' and the jet apex, $r_{\text{apex}}$. We found that $r_{\text{apex}}$, was $\leq 6.0$\,pc upstream of the mm-wave ``core'' and the most upstream site of $\gamma$-ray emission. \item From SSA calculations, we found magnetic field strengths $B_{\text{SSA}} \geq$ 1.6\,G in the ``core'' and $B_{\text{SSA}} \lesssim$ 0.4\,G in the stationary feature. From equipartition, we estimate $B_{\text{equi}} \geq$ 1.2\,G in the ``core'' and $B_{\text{SSA}} \lesssim$ 0.3\,G in the stationary feature. The results from either method are consistent to within $\sim$20\%. \item We also extrapolate estimates of the magnetic field strength to within the BLR and at the jet apex. Using SSA and assuming a toroidal magnetic field, this yields $B_{\text{BLR}} \lesssim 600$\,G and $B_{\text{apex}} \lesssim 4200$\,G. Using equipartition, this yields $B_{\text{BLR}} \lesssim 500$\,G and $B_{\text{apex}} \lesssim 3300$\,G. \end{enumerate} VLBI at 3\,mm currently lacks sensitivity and cadence, although the recently available 2\,Gbps recording modes and the possible addition of phased ALMA to the array will improve the situation \citep{mmvlbi_whitepaper}. Unfortunately with only 6 month intervals, structural changes may be missed, though high cadence monitoring could performed with the eight 3\,mm equipped stations of the VLBA but without trans-Atlantic baselines. In the future, it would be highly desirable to have monitoring at 3\,mm (and 1\,mm) at least as frequent as 43\,GHz monitoring, but with the longest baselines. In a future paper, we will expand this analysis to other blazars observed at 3\,mm to further investigate the physical conditions and dynamics of jets. We will also include polarisation observations, which may prove important in testing the location of $\gamma$-ray emission. \\

1607.00725

1607

1607.07333_arXiv.txt

The gravitational influence of a second satellite on the rotation of an oblate moon is numerically examined. A simplified model, assuming the axis of rotation perpendicular to the (Keplerian) orbit plane, is derived. The differences between the two models, i.e. in the absence and presence of the second satellite, are investigated via bifurcation diagrams and by evolving compact sets of initial conditions in the phase space. It turns out that the presence of another satellite causes some trajectories, that were regular in its absence, to become chaotic. Moreover, the highly structured picture revealed by the bifurcation diagrams in dependence on the eccentricity of the oblate body's orbit is destroyed when the gravitational influence is included, and the periodicities and critical curves are destroyed as well. For demonstrative purposes, focus is laid on parameters of the Saturn-Titan-Hyperion system, and on oblate satellites on low-eccentric orbits, i.e. $e\approx 0.005$.

Saturn's seventh moon, Hyperion (also known as Saturn VII), was discovered in the XIX century by \citet{bond} and \citet{lassel}, but only due to Voyager~2 \citep{smith} and Cassini \citep{thomas10} missions it became apparent that it is the biggest known highly aspherical celestial body in the Solar System, with a highly elongated shape and dimensions $360\times 266\times 205$ km. Since the rotational state of Hyperion was predicted to remain in the chaotic zone \citep{wisdom} based on the spin-orbit coupling theory \citep{goldreich}, further analyses and observations, regarding Hyperion as well as other Solar System satellites, were conducted on a regular basis. Hyperion's long-term observations were carried out twice in the post Voyager~2 era. In 1987, \citet{klav,klav2} performed photometric $R$ band observations over a timespan of more than 50 days, resulting in 38 high-quality data points. In 1999 and 2000, \citet{devyatkin} conducted $C$ (integral), $B$, $V$ and $R$ band observations. The objective of both analyses was to determine whether Hyperion's rotation is chaotic and to fit a solution of the equation of motion to the observations. To the best of the author's knowledge \citetext{Melnikov, priv.\ comm.} there were no other long-term observations that resulted in a lightcurve allowing the determination of Hyperion's rotational state (see also \citealt{strugnell} and \citealt{dourneau} for a list of earlier observations). Although, shortly after the Cassini 2005 passage a ground-based $BVR$ photometry was conducted \citep{hicks}, resulting in 6 nights of measurements (and additional 3 nights of $R$ photometry alone) over a month-long period. Unfortunately, this data was greatly undersampled and period fitting procedures yielded several plausible solutions. The theoretical and numerical treatment of the rotational dynamics of an oblate satellite have been performed widely. After the seminal paper of \citet{wisdom}, \citet{boyd} applied the method of close returns to a sparse and short-term simulated observations of Hyperion's lightcurve. \citet{black} performed numerical experiments using the full set of Euler equations to model long-term dynamical evolution. \citet{beletskii} considered a number of models, including the gravitational, magnetic and tidal moments as well as rotation in gravitational field of two centers. A model with a tidal torque was examined analytically using Melnikov's integrals and assymptotic methods \citep{khan}. The stability of resonances with application to the Solar System satellites was inferred based on a series expansion of the terms in the equation of rotational motion \citep{celletti1,celletti2}. The Lyapunov exponents and spectra were exhaustively examined for a number of satellites\footnote{In particular, Lyapunov times for Hyperion ranged from $1.5\times T$ to $7\times T$, where $T=21.28\,{\rm d}$ is the orbital period.} \citep{shevchenko1,shevchenko,kouprianov1,kouprianov2}. A model of an oblate satellite with dissipation was used to examine the basins of attraction in case of low eccentricities, especially with application to the Moon \citep{celletti3}. The dynamical stability was examined for all known satellites by \citet{melnikov}. Again the dynamical modeling using the full Euler equations was conducted by \citet{harbison}, who also analyzed the moments of inertia in light of the precessional period. Finally, \citet{tarnopolski} argued that in order to extract a maximal Lyapunov exponent from the photometric lightcurve of Hyperion, at least one year of dense data is required. The orbital dynamics of Hyperion in the Saturn-Titan-Hyperion system (see Table~\ref{tbl1} for some physical parameters) have been exhaustively examined due to the interesting 4:3 mean motion resonance between Hyperion and Titan \citep{peale,taylor2,stellmacher,rein}. While the impact of Titan's gravitation on Hyperion's orbit has been established \citep{taylor} and the stability of the resonance has been considered in great detail \citep{colombo,bevilacqua}, introduction of the gravitational impact of a secondary body on the rotation of an oblate satellite has been done before for nearly spherical bodies such as Venus \citep{beletskii2} or low-eccentric orbits in the Pluto-Charon system \citep{correia}. Herein, numerical integrations will be performed within the chaotic zone of the Saturn-Titan-Hyperion system with parameters $\omega^2$ and $e$ such that the perturbation techniques are not valid \citep{maciejewski}, which to the best of the author's knowlege has not yet been examined and hence is the aim of this work, which is general enough to be applicable to moons other than Hyperion. To focus attention, throughout the analysis the parameters are set to those of the Saturn-Titan-Hyperion system unless otherwise stated, but low-eccentricity and low-oblateness cases are also investigated for comparison. \begin{table}[h] \small \caption{Physical parameters of the Saturn-Titan-Hyperion system.} \label{tbl1} \centering \begin{tabular}{@{}cccc@{}} \hline Parameter & Symbol & Value & Reference \\ \hline Saturn's mass & $M$ & $5.68\cdot 10^{26}\,{\rm kg}$ & \citet{jacobson} \\ Titan's mass & $m_1$ & $1.35\cdot 10^{23}\,{\rm kg}$ & \citet{jacobson} \\ & $m_1/M$ & $2.4\cdot 10^{-4}$ & \\ Hyperion's major semi-axis & $a$ & 1 429 600 km & \citet{seidelmann,thomas07} \\ Titan's major semi-axis & $a_0$ & 1 221 865 km & \url{http://ssd.jpl.nasa.gov/?sat_elem} \\ & $a_0/a$ & 0.855 & \\ Hyperion's oblateness & $\omega^2$ & 0.79 & \citet{wisdom} \\ Hyperion's eccentricity & $e$ & 0.1 & \citet{wisdom} \\ Hyperion's orbital period & $T$ & 21.3 d & \citet{thomas07} \\ \hline \end{tabular} \end{table} This paper is organized in the following manner. In Sect.~\ref{model} the rotational models in case of the absence and presence of a second satellite's gravitation are derived. In Sect.~\ref{phase} the phase space is briefly described. Section~\ref{meth} presents the methods used: the correlation dimension and its benchmark testing, and the bifurcation diagrams. The results are presented in Sect.~\ref{res}, which is followed by discussion and conclusions gathered in Sect.~\ref{disc}. The computer algebra system \textsc{mathematica}\textsuperscript{\textregistered} is applied throughout this paper.

\label{disc} The aim of this paper was to investigate how does the gravitational interaction with a second satellite influence the rotational dynamics of an oblate moon. A simplified model was designed, resulting in the equation on motion given in Eq.~(\ref{eq21}), being basically a perturbation of the well known Eq.~(\ref{eq11}). The derived equation of motion introduces a third parameter, the mass ratio $m_1/M$, additional to te oblateness $\omega^2$ and eccentricity $e$. To allow comparison, two sets of ICs distributed uniformly in a $0.1\times 0.1$ square in the phase space were evolved, in case of the absence and presence of the additional source of gravitation. In case of the set IC1 [centered at $(\pi/2,0.55)$] the difference between the two models is qualitative in nature: when the second satellite was absent all trajectories were quasiperiodic (first row in Fig.~\ref{serieplot}), as indicated by the $d_C=1$ in Fig.~\ref{fig7}(a). Interestingly, when its presence was taken into account, one could observe leaking of the orbits into the chaotic sea (second row in Fig.~\ref{serieplot}). This phenomenon manifests itself also through a higher $d_C$ attained in Fig.~\ref{fig7}(b). Hence, it turns out that an additional satellite has the ability to change quasiperiodic orbits into chaotic ones, i.e. it enlarges the chaotic domain. On the other hand, when the set IC2, located in the center of the chaotic region [an $0.1\times 0.1$ square centered at $(\pi/2,1.5)$], was considered, no long term (assymptotic) differences could be observed (third and fourth rows in Fig.~\ref{serieplot}), and the correlation dimension for both models reached a plateau at $d_C\approx 1.75<2$ [Fig.~\ref{fig7}(c) and (d)], likely due to clustering. However, this is not that surprising, given that the gravitational influence under investigation was three orders of magnitude smaller than the planet's, and that the rotational model in absence of the second satellite is dominated by the chaotic zone, hence it would be highly unexpected for it to have the ability to change chaotic motion into a regular one. The bifurcation diagrams, especially interesting when $m_1/M$ and $\omega^2$ were fixed and $e$ was varied, when small values of the eccentricity (i.e., $e<0.03$) are considered lead to a conclusion that the regular and highly structured picture [Fig.~\ref{out1}(d)] becomes much more messy, and the transition to chaos occurs for smaller eccentricities than in the case when additional gravitation source is neglected [Fig.~\ref{out2}(d)]. This is consistent with the results of the other method (i.e., evolving the sets IC1 and IC2) in the sense that the second satellite changes regular motion into chaotic. The differences in case when $\omega^2$ was varied was not that much remarkable [Fig.~\ref{out1}(a) and \ref{out2}(a)], and both models lead to chaotic motion when larger $e$ are considered [Fig.~\ref{out1}(c) and \ref{out2}(c)]. The bifurcation diagram in dependence on the ratio $m_1/M$ was mostly structureless [Fig.~\ref{out2}(b)]. Eventually, the destruction of regular rotation caused by the second satellite might be ascribed to the destruction of the invariant tori (\citealt{tabor}; see also \citealt{celletti1} and references therein). Finally, when $e$ was set to the mean eccentricity of all Solar System satellites (i.e., $e=0.005$), the highly structured bifurcation diagram displayed in Fig.~\ref{out3} also got destroyed and became much more tangled, as shown in Fig.~\ref{out4}. To conclude, the derived simplified model of a second satellite's influence on rotational dynamics of an oblate satellite implies that: \begin{enumerate} \item the additional source of gravitation can change some regular orbits into chaotic ones, and \item destroys the regularity, particularly the periodicities and critical curves, in the bifurcation diagram for small eccentricities $e<0.03$. \end{enumerate}

1607.07333

1607

1607.07619_arXiv.txt

In this paper, which is the first in a series of papers associated with cataclysmic variables and related objects, we introduce the CATUABA code, a numerical machinery written for analysis of the MOCCA simulations, and show some first results by investigating the present-day population of cataclysmic variables in globular clusters. Emphasis was given on their properties and the observational selection effects when observing and detecting them. In this work we analysed in this work six models, including three with Kroupa distributions of the initial binaries. We found that for models with Kroupa initial distributions, considering the standard value of the efficiency of the common envelope phase adopted in BSE, no single cataclysmic variable was formed only via binary stellar evolution, i.e. in order to form them, strong dynamical interactions have to take place. We show and explain why this is inconsistent with observational and theoretical results. Our results indicate that the population of cataclysmic variables in globular clusters is, mainly, in the last stage of their evolution and observational selection effects can drastically change the expected number of observed cataclysmic variables. We show that the probability of observing them during the outbursts is extremely small and conclude that the best way of looking for cataclysmic variables in globular clusters is by searching for variabilities during quiescence, instead of during outbursts. For that, one would need a very deep observation which could reach magnitudes $\gtrsim 27$ mag. Finally, we argue that cataclysmic variables in globular clusters are not necessarily magnetic.

\label{introduction} The study of star clusters plays an important role in our understanding of the Universe since these systems are natural laboratories for testing theories of stellar dynamics and evolution. Particularly, globular clusters (GCs) are one of the most important objects for studying the formation and the physical nature of exotic objects such as X-ray binaries, degenerate binaries, black holes and blue straggler stars, which in turn provides basic information and tools that can help us to understand the formation and evolution processes of star clusters, galaxies and, in general, the young Universe. Among the most interesting objects in GCs are the cataclysmic variables (CVs) that are interacting binaries composed of a white dwarf (WD) that accretes matter stably from a main-sequence (MS) star -- or, in the last stage of their evolution, a brown dwarf (BD) -- \citep[BD; e.g.][for a comprehensive review]{Knigge_2011_OK}. CVs are subdivided according to their photometric behaviour as well as the magnetic field strength of the WD, being, mainly, magnetic CVs (where the accretion is partially or directly via magnetic field lines) and non-magnetic CVs (where the accretion is via an accretion disc). Among the non-magnetic CVs, the most prominent subgroup is that composed of dwarf novae (DNe) which exhibit repetitive outbursts due to the thermal instability in the accretion disc \citep[e.g.][for a review]{Lasota_2001}. \subsection{CV formation} \label{cv_formation} CVs are believed to form from an MS-MS binary that undergoes a common envelope phase (CEP) when the more massive MS evolves to a giant \citep{Paczynski_1976}. In such CEP, the dense stellar giant core and the MS spiral towards each other with the expansion and loss of the common envelope. Most of the angular momentum is lost with the envelope which leads to an orbital period orders of magnitude shorter \citep[e.g.][]{Webbink_1984}. After the CEP, a pre-CV is formed in a detached WD-MS binary. Because of angular momentum loss (see below), the separation between the stars decreases up to the formation of the CV, when the MS starts filling its Roche lobe. In order to form a CV through the above-mentioned scenario, the initial MS-MS binary should approximately have the following properties: (i) the more massive star has $M \lesssim \; 10 \; {\rm M}_\odot$; (ii) the less massive star is a low-mass MS; (iii) the mass ratio is $\lesssim$ 0.25; (iv) the separation has to be sufficient to allow the primary to expand to a point where it could form a degenerate core (pre-WD) and to permit the contact (the CEP). The reasons for these conditions come almost naturally. The mass of the more massive star has to be in the acceptable range to form a WD. This justifies the first condition. Besides, the mass ratio after the CEP has to be $\lesssim$ 1 \citep[e.g.][]{Hellier_2001}. Otherwise, the mass transfer would be unstable and such instability would precipitate further mass transfer and then a merger would occur. This is because the mass transfer rate becomes very large and the WD cannot steadily burn the accreted material, then it swells up to become a giant, producing a common envelope binary and a merger of the stars. The WD mass cannot be greater than $\sim 1.44 \; {\rm M}_\odot$; thus, in order to have the mass ratio less than one, the secondary mass cannot be larger than the WD mass, after the CEP. This justifies the second condition. About the mass ratio, if the initial mass ratio is great (q $\sim 1$), then both stars will have similar evolutionary time-scales. In this way, the stars would become WDs at roughly the same time. For a CV however, one star has to be a WD and the other one an MS. This explains the reason for the third condition. Finally, if the separation was too large, the more massive star could not fill its Roche lobe. While if it was too small, the Roche lobe overfilling could lead to a merger. Then, the separation has to be ideal to allow for both the formation of a WD-like core and the close post common envelope binary. \subsection{CV evolution} \label{cv_evolution} Non-magnetic CVs are usually separated in the following way, with respect to the orbital period: (i) if the donor is an MS and $P_{\rm orb} \gtrsim $ 3 h, then it is called a long-period CV; (ii) if the donor is an MS and $P_{\rm min} \lesssim P_{\rm orb} \lesssim $ 2 h, then it is called a short-period CV; and (iii) if the donor is a BD, it is called a period bouncer CV (in this case, $ P_{\rm orb} \gtrsim P_{\rm min}$). Besides, in CVs (especially non-magnetic ones), angular momentum loss is the driving mechanism for their long-term evolution. In the so-called standard model of CV evolution, the dominant angular momentum loss mechanism in long-period CVs is magnetic braking \citep{Rappaport_1982}, whereas in short-period and period bouncer CVs the driving mechanism is associated with the emission of gravitational radiation \citep{Paczynski_1967}. Basically, two important features observed in non-magnetic CVs need to be explained by the standard model. First, the absence of systems in the range of 2 h $\lesssim P_{\rm orb} \lesssim $ 3 h, known as the period gap, \citep[e.g.][and references therein]{Zorotovic_2016} and, secondly, the existence of a period minimum $P_{\rm min} \approx$ 82 min \citep[e.g.][]{Gansicke_2009}. The standard model reasonably fulfils its role in explaining the observational properties of the CVs. Briefly, the standard model can be summarized as follows: after the birth of the CV, it will evolve towards short periods due to angular momentum loss. When it reaches $P_{\rm orb} \sim$ 3 h (upper edge of the period gap), the donor becomes fully convective and \citet{Rappaport_1983} proposed that at this point the magnetic braking turns off or becomes less efficient (disrupted magnetic braking scenario). This results in a decrease of the mass transfer rate, which allows the donor star to reestablish its equilibrium and to stop filling its Roche lobe. Then, the system becomes detached since the mass transfer stops. However, such a detached WD-MS binary keeps evolving towards shorter periods due to gravitational radiation. When $P_{\rm orb} \sim$ 2 h (inner edge of the period gap), the Roche lobe has shrunk enough to restart mass transfer and the system becomes a CV again. After that, at some point during its evolution ($P_{\rm orb} \sim P_{\rm min}$), the mass loss rate from the secondary drives it increasingly out of thermal equilibrium until the thermal time-scale exceeds the mass loss time-scale and it expands in response to the mass loss, thus, increasing $P_{\rm orb}$. Consequently, a large number of CVs are expected to be near the period minimum (known as the period spike) or be in post-period minimum phase -- indeed, the abundance ratio for long-period CVs, short-period CVs and period bouncers, respectively, is roughly 1:30:70 \citep[e.g.][]{Howell_1997}. \subsection{CV observation} \label{cv_observation} Since the birth of interest in CVs \citep[e.g.] [for a historical review]{Warner_1995_OK}, several breakthroughs have been taking place in the field, especially due to the Sloan Digital Sky Survey which has provided a reasonable sample that reaches deeper magnitudes and which is capable of recognizing very faint CVs near and beyond the period minimum ($P_{\rm min}$). Such breakthroughs include: the confirmation of the disrupted angular momentum loss at the period gap \citep{Zorotovic_2016}; the discovery of period bouncers with BD donors \citep{Littlefair_2006}; the discovery of the period spike around $P_{\rm min}$ \citep{Gansicke_2009}; among others. They allowed the community to considerably improve the observational data to confront theoretical predictions which, in turn, have led to a stronger corroboration between theory and observation. All the previous discussion, theoretical and observational, are mainly concerned with CVs in the field. For CVs in GCs, the same is not always true because of the influence of dynamical interactions, the ages and distances of the GCs, and the corresponding observational selection effects. Some observational efforts have also been made regarding CVs in GCs, especially a search for an optical counterpart of {\it Chandra} X-ray observations \citep[e.g.][for a review on CVs in GCs]{Knigge_2012MMSAI}. In general, there are four main approaches that have been used for detecting CVs in GCs. In what follows, we will describe them briefly including also the few important works associated with them. \subsubsection{Variability during outbursts} \label{cv_observation_variability} In all possible sorts of variability related to CVs, the most explored is that regarding the DN outbursts which last from days to tens of days and result in an increase of luminosity by roughly 2 -- 5 mag. The first major investigation of CVs in GCs was done by \citet{Shara_1996}, who analysed 12 epochs of {\it Hubble Space Telescope (HST)} observations of the GC 47 Tuc and recognized only one DN and one DN candidate. Other DNe have been discovered through variability during outbursts after that in different clusters \citep[e.g.][]{Shara_2005,Knigge_2003,Servillat_2011}. In a survey-like search for DNe, \citet{Pietrukowicz_2008} analysed 16 Galactic GCs and yielded two new certain DNe: M55-CV1 and M22-CV2. They found that the total number of known DNe in the Galactic GCs is 12 DNe, distributed among seven clusters. Some comments are needed at this point. As \citet[][Section 5.5]{Knigge_2012MMSAI} has already pointed out, both \citet{Shara_1996} and \citet{Pietrukowicz_2008} concluded that DNe are rare in GCs based on the properties of observed CVs in the Galactic field, which seems to be a biased sample of the real population of CVs in the field. If most CVs in the field are, in fact, period bouncers, then the observed CV population in the field (especially the bright ones) is not representative of the real population of CVs in the Galactic field. This is mainly due to the very small duty cycle (fraction of the DN cycle that a DN is in outburst) associated with period bouncers. Hence, in this case, most CVs in the field are unobservable at any given time, and a significant population of hidden CVs exists. Thus, the conclusion that DNe are rare in GCs is not necessarily correct, since it could just be that they are hard and/or unlikely to observe, since there are more period bouncer CVs than originally expected \citep{Knigge_2012MMSAI}. Therefore, it turns out that identifying CVs through their variability during outbursts is unlikely to reveal the intrinsic population of CVs in GCs since one should be very lucky to detect the outbursts in a sequence of epochs from {\it HST}, given the extremely small duty cycle associated with the CVs in GCs. \subsubsection{H$\alpha$ imaging} \label{cv_observation_haplha} Another possible way to detect CV candidates is using H$\alpha$ imaging \citep[e.g.][]{Cool_1995}, since systems that exhibit an excess in H$\alpha$ show evidence of variability. This technique is generally used to study the counterparts of X-ray sources and has revealed few CVs in some GCs \citep[e.g.][]{Grindlay_2006,Pietrukowicz_2009}. However, doubt remains about the deepness of observations using these techniques, i.e. if they are able to detect the faint population of CVs in GCs. So far, \citet{Cohn_2010} have reached magnitudes as deep as 28 for H$\alpha$ and 26 from optical observations of the {\it HST}. Their study seems to be the least affected by this kind of bias and will be used in this paper as the object for comparison with our results. \subsubsection{FUV band with {\it HST}} \label{cv_observation_fuv} Another way to detect potential CVs is through their colours. CVs tend to be bluer due to accretion processes. In fact, the energy released from this mechanism makes the region close to the WD hotter, which in turn, makes the CVs bluer than typical stars in GCs. This implies that looking for them in the far-ultraviolet (FUV) band with {\it HST} is a good way to find CV candidates \citep[e.g.][]{Dieball_2010}. Especially because most MSs in GCs emit in infrared which nulls the problem with crowding. \citet{Dieball_2010} carried out a detailed search in the core of M80 and found few candidates. However, due to their limiting magnitude, they could only detect bright CV candidates. In this way, the detection of the faint CVs using the {\it HST}'s FUV detectors might also be problematic. \subsubsection{X-rays} \label{cv_observation_fuv} The high resolution that has been achieved with {\it Chandra} allows us to, in fact, reach binaries with compact objects in GCs, especially in their cores. With regards to CVs, many GCs have been studied with {\it Chandra} down to $\sim 10^{32}$ erg s$^{-1}$ \citep[e.g.][]{Pooley_2010}. Additionally, the identification of optical counterparts with deep {\it HST} imaging has allowed for the recognition of many CV candidates \citep[e.g.][]{Bassa_2004,Huang_2010,Cohn_2010}, although the number of such candidates is far from the predicted number of CVs in the observed clusters. Finally, it is valid to note that below $\sim 10^{32}$ erg s$^{-1}$, a large fraction of X-ray sources do not have secure optical counterparts. Below this value, the sources can be chromospherically active binaries (or near-contact binaries of MSs), CVs, foreground and background objects, quiescent low-mass X-ray binaries, millisecond pulsars or black hole X-ray binaries. Any conclusions drawn from a comparison between the results of our simulations and observations of CVs with small X-ray luminosities should be taken with a grain of salt. This is because the observational sample can be regarded as something of an upper limit, due to an increased probability of contamination from active binaries, chromospherically active stars, accreting neutron stars and black holes, etc. \subsubsection{Classical novae} \label{cv_observation_cn} On the subject of classical novae (CVs with high mass transfer rates and stable and hot discs), it is worth mentioning some observational evidence of different nova rates with respect to the Galactic field. \citet{Curtin_2015} detected novae in a survey of GCs in three Virgo elliptical galaxies (M87, M49 and M84). Such a survey should not detect any novae if there were no enhancement of the nova rate due to dynamics. A similar result was reached by \citet{Shara_2004} while investigating one GC of M87. They concluded that classical nova eruptions in GCs are up to 100 times more common than current detections in the Milky Way suggest. This implies that dynamics are extremely important in enhancing the detection rate of novae in GCs. \subsubsection{What is the lesson from the observations?} \label{lesson} Given the crowding of GCs and the faintness of the intrinsic population of CVs, confirming spectroscopically the many CV candidates that have been observed as real CVs is challenging \citep[e.g.][]{Knigge_2003,Thomson_2012}. On the other hand, the usage of a combination of different techniques (H$\alpha$ and FUV imaging, X-ray, colour, etc.) can provide almost guarantees the confirmation of CVs, especially for DNe. For instance, \citet{Cohn_2010} used H$\alpha$ imaging and colours to infer that some {\it Chandra} X-ray sources are CVs. As it seems, a combination of techniques can provide us with the potential number of CVs in GCs. The only problem is whether or not we can reach faint CVs, given the observational limitations and biases of each technique when combined together. This indicates that the biases and observational limitations can lead us to incorrect impressions about the nature of CVs in GCs. \subsection{Nature of CVs in GCs} \label{cv_nature} In the most recent survey-like search for DNe, \citet{Pietrukowicz_2008} conclude that `ordinary DNe are indeed very rare in GC'. Then, either the predicted number of CVs is not correct, or the predicted CVs would be non-DNe. Nevertheless, such observational findings do not corroborate theoretical predictions. First, theoretical works predict that most CVs should be DNe \citep[e.g.][]{Knigge_2011_OK}. Secondly, some previous numerical investigations regarding CVs in evolving GCs estimate around 100-200 CVs in massive GCs. In the most recent study on the matter, \citet{Ivanova_2006} predicted that around 200 CVs would be present in a typical massive GC. Besides, such CVs would have different properties from the CVs in the field. For example, only $\sim$ 25 per cent of CVs were formed in binaries that would become CVs in the field. Also, approximately 60 per cent of the CVs did not form via a CEP. This corresponds to a rather strong inconsistency between theory and observation and the most popular hypothesis that attempts to explain the so-called absence of DNe in GCs is based on the mass transfer rate and the WD magnetic field. \citet{Dobrotka_2006} proposed, using the disc instability model (DIM), that a low mass transfer rate combined with a moderately strong WD magnetic field can disrupt the inner part of the disc, preventing, in turn, an outburst in such CVs. In the mid-1990s, the community started to think that CVs in GCs tend to be magnetic due to the work by \citet{Grindlay_1995} who analysed three CVs in NGC 6397 and found the existence of He II line in such CVs. Also, \citet{Edmonds_2003a,Edmonds_2003b} argued in the same direction in their studies of 47 Tuc. The big problem with this argument is that not only magnetic CVs exhibit the He II line in their spectra, but also other types of CVs \citep{Knigge_2012MMSAI}. For instance, \citet{Shara_2005} found DN-like eruptions in CV2 and CV3 in NGC 6397 which exhibited helium emission in their spectra. Therefore, it seems that such evidence is not strong enough to claim that CVs in GCs are, principally, magnetic ones. The main reason for this suspicion is the attempt to explain the discrepancies between observed CVs in GCs and those in the field. This is because a WD magnetic field can prevent instability in the disc and, in turn, suppress the occurrence of outbursts \citep{Dobrotka_2006}. Besides, it could explain to some extent, the unique X-ray and optical properties found for the CVs in GCs. However, the CV samples in GCs tend to be X-ray selected \citep{Heinke_2008} which, in turn, favours the detection of magnetic CVs (brighter in X-ray than the non-magnetic counterpart). Unfortunately, few investigations went deep enough to detect the non-magnetic CVs ($\lesssim 10^{30}$ erg s$^{-1}$) in the X-ray and more efforts should be put in this direction. Another point in favour of the idea that CVs in GCs are magnetic comes from the fact that magnetic CVs are usually associated with massive WDs \citep{Ivanova_2006}. In fact, in GCs, the dynamically formed CVs tend to have higher WD masses which, in turn, favours the hypothesis of magnetic CVs. Above all, such a hypothesis is not well established and can be contested. As has already been mentioned, most CVs should be DNe \citep{Knigge_2011_OK}. Besides, not only magnetic CVs produce the above-mentioned He II emission. Over and above, many optical counterparts of X-ray sources have been recognized as reasonable CV candidates in some GCs \citep[e.g.][]{Cohn_2010}, although such numbers are still far from the predicted number of CVs. \subsection{Structure of the paper} \label{paper_structure} From the last subsection, we can say that the `CV problem in GCs' is not yet well understood and solved. In order to contribute to such a discussion, this is the first paper of a series that concentrates on CVs and related objects (such as AM CVn and symbiotic stars) in evolving GCs that attempt to correlate CV properties (and also AM CVn and symbiotic star properties) with cluster initial and present-day parameters. The main objective of this first paper is to introduce the CATUABA code (Section \ref{catuaba}) that will be used in this series of papers to derive properties of CVs and related objects from MOCCA simulations. In order to test its efficiency and coherence, we concentrate on only six models with different initial conditions and different properties at the present-day (described in Section \ref{models}). The models were simulated by the MOCCA code (Section \ref{mocca}) that simulates real GCs on a time-scale of one day. Its speed allows for the simulation of hundreds of models in a short time and, in turn, permits a very detailed statistical analysis of particular objects (like CVs) and their correlations with the cluster parameters. The aim of this series of papers is to analyse the MOCCA database with respect to CVs and related objects. For convenience, we decided to separate the results of this initial work into two separate papers. In this paper we will concentrate mainly on the present-day (considered here as 12 Gyr) population of CVs present in the analysed models. In the next, we will discuss mainly the CV progenitors and the main formation channels as well as their properties at the moment they are formed and the subsequent evolution up to present day. We will also deal with more general issues like unstable systems, escaping binaries that become CVs, and so on. In Section \ref{mocca}, we describe the MOCCA and BSE codes. We also make some comments with regards to the comparison between MOCCA and N-body codes. We describe the CATUABA code in Section \ref{catuaba}, and end the Section by summarizing its main features. The models used in this first work will be described in Section \ref{models}, and in Section \ref{results}, the main results of the preliminary investigation are presented. We show results associated with the clusters and their present-day populations (PDPs) of CVs as well as results related to observational selection effects when searching for CVs in GCs. We also address some connections between our results and observations and, also, between our investigation and previous studies. We conclude and discuss the main implications of these first results in Section \ref{conclusion}. Finally, throughout this paper, we use some new abbreviations, and for convenience, in Appendix \ref{ap1}, we clearly define all abbreviations in order to allow the reader to consult them if necessary.

\label{conclusion} So far, we have exposed in this paper part of the first results obtained from the analysis of six models simulated by MOCCA using the CATUABA code as well as description of the code itself. The PDP of CVs in GC are in the last stage of their evolution! This, maybe, is the leading characteristic of the CVs in GCs. In addition, if the WD magnetic field is irrelevant, they all are DNe. In order to check if one could `easily' observe the DNe with typical limiting magnitudes in previous surveys, a simple (and rather ideal) probability law was derived and computed, whose results are displayed in Fig. \ref{Fig06}. It shows clearly that observational selection effects are hiding the real population of CVs in GCs. This can also be thought of with respect to the CVs in the field \citep{Pretorius_2007}. In order to observe the population of CVs in GCs, one should be able to perform deeper observations in such a way that the search would be during quiescence. Something that was already noticed by \citet{Knigge_2012MMSAI} and, to some extent, done by \citet{Cohn_2010}. Interestingly, some alternatives exist in order to look for CVs during quiescence. For instance, \citet{Cohn_2010} used H${\rm \alpha}$ imaging (which allows a search for short-period variability) in combination with optical observations of {\it Chandra} X-ray sources. Other ways that might help are FUV with {\it HST} \citep{Knigge_2003} and quiescent negative superhump \citep{Olech_2007}. In any event, we defend here that the CVs in GCs are not necessarily magnetic. The arguments for magnetic CVs seem to be biased in addition to neglecting some other issues, like observational selection effects and incompleteness in surveys. As has already pointed out by \citet{Knigge_2012MMSAI}, the natural solution to this problem is a survey for DNe in GCs that guarantees the detection of at least a few WZ Sge systems. More about the CV populations like the progenitor population, the formation-age population, and also the evolutionary properties will be discussed, as was said before, in a separate work. Additionally, the apparent problem entailed with the Kroupa IBP is also left to another paper since more models will be needed and an attempt to overcome such an inconsistency will be discussed. Thus, for this paper, the results displayed in Section \ref{results} can be summarized as follows. (i) GCs with Kroupa IBP have no CVs formed purely via binary stellar evolution, i.e. dynamics have to act in order to produce CVs. The reason can be either the efficiency of the feeding algorithm or the high CEP efficiency adopted in BSE. (ii) The main mechanism in CV formation for the Kroupa models is exchange ($\sim \; 2/3$), which is not true for other models (Standard models) in which the main mechanism is the binary stellar evolution. (iii) Almost all CVs in the present-day clusters are formed below the gap. Additionally, most of the CVs at 12 Gyr ($\sim$ 87 per cent) are period bouncers with BD as donors, having periods between 1.5 and 3h. (iv) We obtained, as in previous simulations, a population of CVs with WDs with small masses ($\sim$ 0.3 ${\rm M_\odot}$). Moreover, in our case, the massive WDs ($\sim$ 0.7 ${\rm M_\odot}$) correspond mainly to CVs formed because of strong dynamical interactions (mainly exchanges), which is in good agreement with previous works \citep[e.g.][]{Ivanova_2006}. (v) All CVs are DNe if there is no strong WD magnetic field such that it can disrupt the inner part of the disc. (vi) The duration of the outburst varies from 1-10 days and the recurrence time (duration of the quiescent phase) varies from 10-10000 days. This implies that the probability of detecting most of the CVs during outburst is extremely low, even considering an ideal situation in which all nights -- during the DN cycle -- are observed. (vii) During quiescence, the eclipsing CVs would be detected in a very deep observation (apparent visual magnitude $\gtrsim 27$ ) in a very close cluster ($\lesssim 5$ kpc). (viii) GCs are old objects which implies that the population of CV tends to be older than the observed population in the field. Thus, at the present-day (12 Gyr), the population of CVs is dominated by low-mass donors. In other words, the properties of the population of CV changes with time through the cluster evolution and this has to be taken into account for meaningful comparisons between observations and predictions. (ix) in order to solve the problem regarding the nature of CVs in GCs, additional effort should be put into the optical identification of faint {\it Chandra} X-ray sources ($\lesssim 10^{30}$ erg s$^{-1}$) and the utilization of other techniques (e.g. H$\alpha$ and FUV imaging with {\it HST}, quiescent negative superhumps) combined together with the aim of detecting almost guaranteed DNe, specially because any conclusions drawn from a comparison between the results of our simulations and observations of CVs with small X-ray luminosities should be taken with a grain of salt, since the observational sample can be regarded as something of an upper limit, due to an increased probability of contamination from active binaries, chromospherically active stars, accreting NSs and BHs, etc. (x) the best region inside GCs to search for CVs should be between the core and half-mass radii. Finally, some comments are needed concerning the Monte Carlo approach and the population synthesis code utilized in this work. In Section \ref{nbody_codes} we briefly described some comparisons made between MOCCA and NBODY6 that have led to good agreement between the two. Therefore, we think the approach chosen in this work is reasonable and efficient since the MOCCA code is significantly faster than the most advanced version of NBODY6++GPU. Even then, the main limitation turns out to be related to the binary stellar evolution code. Some limitations with regard to the CV evolution were commented on in Section \ref{bse_code}, like the absence of expansion of the donor star for long-period CVs and the old scaling for the angular momentum losses above and below the gap. Nevertheless, from a statistical point of view, such improvements should not change the main conclusions drawn in this work.

1607.07619

1607

1607.07934_arXiv.txt

We measure the star formation quenching efficiency and timescale in cluster environments. Our method uses N-body simulations to estimate the probability distribution of possible orbits for a sample of observed SDSS galaxies in and around clusters based on their position and velocity offsets from their host cluster. We study the relationship between their star formation rates and their likely orbital histories via a simple model in which star formation is quenched once a delay time after infall has elapsed. Our orbit library method is designed to isolate the environmental effect on the star formation rate due to a galaxy's present-day host cluster from `pre-processing' in previous group hosts. We find that quenching of satellite galaxies of all stellar masses in our sample ($10^{9}-10^{11.5} \Msun$) by massive ($> 10^{13} \Msun$) clusters is essentially $100$~per~cent efficient. Our fits show that all galaxies quench on their first infall, approximately at or within a Gyr of their first pericentric passage. There is little variation in the onset of quenching from galaxy-to-galaxy: the spread in this time is at most $\sim 2$~Gyr at fixed $M_*$. Higher mass satellites quench earlier, with very little dependence on host cluster mass in the range probed by our sample.

A detailed understanding of the mechanisms that quench star formation in galaxies remains elusive. It now seems clear that quenching is strongly correlated with an `internal' parameter that is closely related to galaxy mass: stellar mass \citep{2003MNRAS.341...33K, 2004ApJ...600..681B}, velocity dispersion \citep{SmiLucHud09c,GraFabSch09}, or structural properties, such as the central stellar surface mass density \citep{CheFabKoo12, FanFabKoo13} or the bulge fraction \citep{OmaBalPog14, 2014MNRAS.441..599B}. The physical cause of this quenching is still not known, although AGN \citep{2004ApJ...600..580G, 2006MNRAS.370..645B, 2006MNRAS.365...11C} and/or mergers \citep{2006ApJS..163....1H} and/or disc instabilities \citep{2009ApJ...703..785D} are often cited. It has become clear that environment also plays a role: once a galaxy falls into a more massive halo (such as a group or cluster) and becomes a satellite, there is an additional probability of quenching over and above the stronger `mass-related' quenching \citep{BalBalNic04, 2008MNRAS.387...79V, 2010ApJ...721..193P}. It is this latter `satellite quenching' that is the subject of this paper. An infalling, actively star-forming satellite galaxy might be affected by its host halo in several ways, as reviewed by \citet{BosGav06}. Ram pressure stripping may remove cold gas from the disc \citep{1972ApJ...176....1G}. It has long been known that cluster galaxies are HI-deficient \citep[e.g.][]{GioHay85}. Furthermore, ram pressure stripping has been observed in individual infalling galaxies in a number of nearby clusters. In particular, in the Coma cluster at least $40$~per~cent of blue galaxies within $500$~kpc of the centre have young stars formed from stripped material visible at ultraviolet wavelengths \citep{2010MNRAS.408.1417S}, which suggests that ram pressure stripping is ubiquitous. However, from these snapshots it is difficult to tell how rapid this process is, and how effective it is overall. A closely-related physical process is `strangulation' \citep{1980ApJ...237..692L, 2000ApJ...540..113B} in which the hot gas halo is stripped by ram pressure, thus removing the source that would otherwise have replenished the cold gas in the disc. Because the cold gas is not immediately affected, the timescale for strangulation should be longer than ram pressure stripping of the cold gas disc. One way to constrain the quenching mechanism(s) is by measuring how effective quenching is, where or when it first occurs in the satellite's orbit, and how long a satellite takes to quench. Environment has a role in regulating star formation on a range of host mass scales \citep[e.g.][]{2010ApJ...721..193P}. In this paper we focus on galaxy clusters. These lend themselves well to our methodology for two reasons. First because their centers and extents, in the sense of both position and velocity, are much better defined than in poorer galaxy groups (though centering becomes easier again for systems such as the Milky Way and its satellites). Second, our choice to study relatively large satellites around massive clusters means we are able to draw our sample of observed galaxies from a large volume, yielding a statistically powerful data set. While clusters host a relatively small fraction of the passive galaxy population -- only about 15~per~cent (2.5~per~cent) of red galaxies with $M_* \geq 10^{9}$\Msun are satellites in $\geq 10^{13}$\Msun ($10^{14}$\Msun) haloes (estimated using the galaxy stellar mass function of red galaxies from \citealp{2012MNRAS.421..621B} and the satellite fraction and host halo mass distribution of \citealp{2008MNRAS.387...79V}) -- they offer a useful proving ground for analysis techniques aimed at constraining the timescale(s) of the quenching process(es) before attempting to tackle the more difficult galaxy group scale. In addition, it may be that a single environmental quenching mechanism is dominant in host haloes of all masses (see for instance \citealp{2008MNRAS.387...79V}, though \citealp{2015arXiv150306803F} argue the opposite). In this case the study of quenching in clusters can directly inform more difficult studies of lower mass hosts. Some early work involved comparing semi-analytic models to observations. In models where the quenching occurs quickly after crossing the host halo's virial radius, too many red satellites were produced. The disagreement suggests that quenching process had to be slow \citep{WeivanYan06a, 2008MNRAS.389.1619F, BalMcGWil09, WeiKauvon09, 2009MNRAS.394.1131K}. \citet{2010MNRAS.409..405H} showed that, while bulge colours do not depend on cluster-centric radius, the colours of discs are redder closer to the cluster centre. These results were modelled by \citet{TarHudBal14}, who found that star formation in discs declines with an exponential timescale of $\sim 3$~Gyr, starting at cluster infall. \citet[hereafter W13]{2013MNRAS.432..336W} concluded that satellites are quenched on timescale of 2-6 Gyr after passing the virial radius of a larger host halo for the first time. W13 obtained this result by measuring quenched fractions of satellites and centrals in low-redshift SDSS groups and clusters, and comparing these data with a satellite quenching model based on a halo infall time distribution from N-body simulations combined with an empirically-calibrated model of the quenched fractions at higher redshifts. Similar results to those of W13 were obtained by \citet{2014MNRAS.444.2938H}. At higher redshifts ($z \sim 1$), \citet{2014MNRAS.438.3070M} found shorter timescales of order $1$~Gyr. An alternative approach is to take advantage of galaxies' positions in the observational projected phase space (PPS) of separation in the plane of the sky and line of sight velocity. \citet{2005MNRAS.356.1327G} showed that, at the same projected radius, galaxies in different phases of their orbits have different kinematics. This was extended by \citet{2013MNRAS.431.2307O} who constructed a subhalo orbit library that allowed them to construct a detailed probabilistic mapping between position in PPS and subhalo infall time. \citet{2011MNRAS.416.2882M} were the first to deproject the PPS to obtain constraints on star formation histories of galaxies falling into larger systems. They studied galaxies with recent (within $1-3$~Gyr) or ongoing star formation, and concluded that star formation is efficiently quenched in a single passage through the cluster. Other authors have used the PPS to understand quenching at high redshift \citep{2014ApJ...796...65M} or the effects of ram pressure stripping \citep{2014MNRAS.438.2186H,2015MNRAS.448.1715J}. The aim of this paper is to study the star-formation rates of galaxies based on their location in PPS, and model these using the orbit libraries of \citet{2013MNRAS.431.2307O}. Whereas \citet{2011MNRAS.416.2882M} used a coarse binning of the populations (`virial', `backsplash', `infalling'), in this paper we use detailed orbit libraries drawn from N-body simulations. Our model consists of two components: (1) an infalling population of galaxies (which are observed predominantly outside the virial radius and are assumed to have some `pre-processed' quenched fraction) and (2) a simple model for quenching in which some fraction of the active infalling galaxies are quenched following a delay $\Delta t$ after passing $2.5\,r_{\rm vir}$. The model then predicts the quenched fraction at any point in PPS. The infalling quenched fraction is fit from the PPS data simultaneously with the free parameters of the model (the efficiency of quenching and the timescale of quenching). This allows us to account for `pre-processing' in a natural way, and hence our results isolate the physical effects of infall of active satellites into their \emph{current} cluster-mass ($\sim 10^{14.5} \Msun$) host haloes. This differs from the approach of W13, in which the quenching timescale refers to the time since infall into \emph{any} halo, and so includes processing in the current host halo plus `pre-processing' in host haloes of lower mass. This paper is structured as follows: in \S\ref{sec-data} we describe our numerical and observed data samples. In \S\ref{sec-method} we describe our models and fitting method. In \S\ref{sec-results} we present the results of fitting our models to the observed data. We discuss our results and compare to other work in \S\ref{sec-conclusions} and summarize in \S\ref{sec-summary}. We assume the same cosmology used in the Bolshoi and Multidark Run 1 simulations with $\Omega_m=0.27$, $\Omega_\Lambda=0.73$, $\Omega_b=0.0469$, $n_s=0.95$, $h_0=0.70$, $\sigma_8=0.82$ \citep{2012MNRAS.423.3018P}.

\label{sec-conclusions} In the results of fitting our model, shown in Fig.~\ref{mstar_trends}, we note the expected trend in $f_{\rm passive,out}$, with higher stellar mass galaxies outside the clusters that are more affected by `internal quenching' having higher $f_{\rm passive,out}$. Interestingly, in all cases we recover values of $f_{\rm passive, in}$ consistent with $1.0$ within the $68$~per~cent confidence interval, and in most cases the best fit value is $\approx 1.0$, suggesting that quenching by clusters is $100$~per~cent efficient\footnote{Indicating that, once quenching by the cluster has had time to operate, $100$~per~cent of satellites have been quenched, but not that the cluster is responsible for quenching $100$~per~cent of the passive galaxies it contains.}, in agreement with the conclusions of \citet{2011MNRAS.416.2882M}. Of course, the \emph{observed} fraction of quenched satellites will always be less than this, even in radial bins closest to the cluster centre, because these bins contain a fraction of satellites falling into the cluster for the first time, which have not yet had time to be quenched, and galaxies `projected into' the cluster. We find a flat or perhaps slightly decreasing trend in $\Delta t$ with increasing stellar mass, and usually small values of $\tau$ (however, in a few cases up to several Gyr, though with $95$~per~cent confidence intervals still consistent with near-zero values). We examine the details of these trends in more detail below. In the upper panel of Fig.~\ref{mstar_trends}, the lower best fit values of $f_{\rm passive,in}$ in the lowest $M_*$ bin stand out as peculiar (also, to a lesser extent, for the lower $M_{\rm host}$ bin in the $9.5<\log_{10}(M_*/\Msun)<10$ bin). While this could be a sign that these lower mass galaxies are more resiliant to quenching, considering other peculiarities in the fits in these bins we cautiously prefer an interpretation where $f_{\rm passive, in}\approx1.0$ in all cases. We first point out that the $68$~per~cent confidence intervals extend up to $1.0$ in all cases. In all cases the marginalized posterior probability distribution for $f_{\rm passive,in}$ (not shown) peaks at $1.0$, but in these peculiar cases the global maximum likelihood is offset from the peak of the marginalized distribution. We believe that this is due to the proximity to the mass resolution limit in the simulations, which causes the satellite haloes contributing orbits for use in this bin to be biased more toward the upper edge of the bin than they would otherwise be. \citet{2013MNRAS.431.2307O} showed that higher mass satellite haloes, which host higher mass galaxies, have orbits with preferentially smaller backsplash distances, which is easily understood as the effect of dynamical friction. If a PDF constructed from a collection of satellite orbits biased toward higher masses is used, when fitting the model, the lower mass galaxies have inferred times since infall that are biased low, driving down the fit quenching timescales ($\Delta t$, $\tau$, or both). This picture seems consistent with the timescales plotted in the lower panel of Fig.~\ref{mstar_trends}, particularly for the higher $M_{\rm host}$ bin, which has a seemingly unrealistic\footnote{Keeping in mind our definition of infall at $2.5\,r_{\rm vir}$} best-fit $\Delta t=0$. This has a knock-on effect on $f_{\rm passive,in}$, driving the best fit to lower values. This situation is exacerbated by the relatively low numbers of observed cluster satellite candidates in these mass bins (see Fig.~\ref{obs_masses}). These difficulties are reflected in the statistical uncertainties derived from the posterior distribution; with the exception of those for $f_{\rm passive,out}$, which is constrained primarily by the properties of interlopers, these are very large. The trends seen in the $\tau$ parameter are also puzzling at first glance. This parameter turns out to be difficult to constrain using our methodology, with $68$~per~cent confidence intervals up to several Gyr wide. Inspecting the marginalized posterior distributions (e.g. Fig.~\ref{banana}), we invariably find a strong degeneracy between $\tau$ and $\Delta t$. This is intuitive, as a rapid transition at a given time is numerically similar to a slightly slower transition that begins slightly earlier. We obtain tighter constraints by considering a representative single timescale $t_{1/2}$. The trends and $68$~per~cent confidence intervals for this parameter combination for the higher (lower) $M_{\rm host}$ bin as a function of $M_*$ are illustrated by the solid (dashed), red (pink) lines and corresponding shaded regions in Fig.~\ref{fig_compare}. The intervals remain large for the subsamples with relatively low observed galaxy counts, but we verify via fits to a Monte Carlo sampling of the quenching timescale distributions that the decreasing trend with increasing $M_*$ is significant at $92$~per~cent ($61$~per~cent) confidence for the lower (higher) $M_{\rm host}$ bin. Further efforts to understand these trends would likely benefit from a more sophisticated model which explicitly models the trends and fits data across the entire $M_*$ and $M_{\rm host}$ range simultaneously. We performed two tests to investigate the effect of changing the information contained in the infall time PDFs. In both cases we used the same subsample used illustratively in Fig.~\ref{banana}. First, we reconstructed our PDFs binning only along the $R$ direction in the $(R,V)$ plane, effectively ignoring the velocity information and emulating the scenario where robust redshifts for cluster members are unavailable. The fit using this modified PDF is broadly similar to the one using the PDF including the velocity information. The preferred $\tau$ drops from $0.60^{+0.45}_{-0.45}$ to $0.00^{+0.15}_{-0.00}$~Gyr, and $\Delta t$ increases from $5.15^{+0.73}_{-0.70}$ to $5.74^{+0.14}_{-0.35}$~Gyr. The values are consistent within the quoted confidence intervals, and we note that, though $\Delta t$ and $\tau$ vary individually, the combined timescale $t_{1/2}$ increases by only $0.18$~Gyr. The statistical uncertainties on all parameters are somewhat narrower when the velocity information is not used, which at first seems surprising. However, the maximum likelihood drops from $-2702$ to $-2721$, a formally very significant ($\sim5.3\sigma$) difference. The narrower confidence intervals are a natural consequence of the poorer fit: the $\chi^2$ is larger, so the change in a parameter required to produce a given change in $\chi^2$ shrinks, apparently leading to narrower confidence intervals, but this is an illusion due to a poorer model fit. The second test we performed was to reconstruct the infall time PDFs by dividing the $(R,V)$ plane into $50$ bins in each direction (our fiducial PDFs use $100 \times 100$ bins). In this case we recover a formally somewhat better fit, with the likelihood increasing from $-2702$ to $-2697$ ($\sim2.2\sigma$ significant). The best-fitting parameters are consistent within the confidence intervals; $\tau$ drops to $0.00^{+0.45}_{-0.00}$~Gyr and $\Delta t$ increases to $5.76^{+0.12}_{-0.36}$~Gyr, again highlighting the degeneracy between the two parameters. For this reason we prefer to focus on the `combined' timescale $t_{1/2}$, but we note that our conclusion that $\tau$ prefers small values $\lesssim 2$~Gyr appears to be robust. \subsection{Comparison with other works}\label{subsec-compare} Our timescales are not directly comparable to many previous studies of satellite quenching for two reasons. First, the radius at which `infall' is defined ($2.5\,r_{\rm vir}$) is larger than most previous studies which typically adopt $1.0\,r_{\rm vir}$ (with varying definitions of `virial'). As noted above, we chose this large radius to avoid the ambiguity of tracking `backsplash' subhaloes which would otherwise exit and re-enter the virial radius. A correction for this difference is relatively straightforward, since the time for a typical subhalo to move from $2.5\,r_{\rm vir}$ to $1.0\,r_{\rm vir}$ is $\sim 3$~Gyr (in detail this depends on which virial definitions are assumed). Second, some previous studies define the time for quenching since the \emph{first} time a subhalo falls into a larger halo of \emph{any} mass. Thus, for example, $30$~per~cent of satellites falling into a $\sim 10^{14} \Msun$ cluster halo had already become satellites of a lower mass group that then fell into the cluster-mass halo. Their quenching time therefore includes the time a satellite spent being `pre-processed'. In contrast, our methodology compares a quenched population ($f_{\rm passive, in}$) with an infalling population that is already `pre-processed' ($f_{\rm passive, out}$) and so isolates the quenching that is due only to falling into the current $\sim 10^{14} \Msun$ host halo. Consequently, due to the different definitions, if the infall radii were the same, our times since infall would always be shorter. \begin{figure*} \leavevmode \epsfxsize=2\columnwidth \epsfbox{figs/fig9} \caption{Comparison of our combined timescale $t_{1/2}$ values (lines) and $68$~per~cent confidence intervals (shaded regions) as a function of $M_*$ with those of W13 (shaded regions represent $68$~per~cent confidence intervals) and \citet[][horizontal errorbars representing the interquartile range of $M_*$ for their sample, vertical error bars representing the uncertainty on satellite quenched fraction from 25 to 55~per~cent]{2014MNRAS.442.1396W}. Host mass ranges or representative values are as shown in the legend. In both cases, their definitions of infall time differ from ours, but we attempt to correct for the differences. See \S\ref{sec-conclusions} for the details of these corrections.\label{fig_compare}} \end{figure*} In Fig.~\ref{fig_compare} we show a comparison of our results with those of W13. We compare our combined timescale $t_{1/2}$ with their $t_Q$ parameter, which is similarly a combination of a delay and a transition timescale (though in W13 the `transition' time refers to the time for an individual galaxy to `fade' from blue to red). In order to compare quantitatively -- W13 uses a very different methodology to ours, but a sample with overlapping mass cuts -- we have attempted to make a correction for the different definitions mentioned above. We correct for the offset between first infall (their preferred definition) and recent infall (which corresponds to crossing $r_{\rm 200b}$), which depends on the host mass\footnote{The difference between first infall and most recent infall also depends on $M_*$, but this is a much weaker effect (Wetzel, private communication) that we neglect here.}, using the data from their fig.~2. We also correct for the median time between crossing $2.5\,r_{\rm vir}$ (i.e. our infall time) and crossing $r_{\rm 200b}$, which is $2.5$~Gyr. We make a further small correction for `ejected' (i.e. `backsplash') haloes using the timescales in \citet{2014MNRAS.439.2687W}. The total offsets we apply to the W13 results for their three host mass bins (low to high) are $2.2$, $1.1$ and $-0.1$~Gyr. The comparison is shown in Fig.~\ref{fig_compare}. We find the same trend of a decreasing quenching timescale with increasing $M_*$, though it appears the slope in our results is somewhat shallower. We also find a much weaker trend than W13, perhaps no trend, with $M_{\rm host}$ in the range probed by our sample. Some of the difference may be explained by the different treatment of `pre-processing'. At the high $M_*$, high $M_{\rm host}$ end, W13 find quenching times that, in our interpretation, correspond to quenching over a Gyr \emph{before} first entering $r_{\rm vir}$. This seems likely to be the signature of pre-processing in another group or cluster. In contrast, in our methodology which treats `pre-processed' galaxies simply as part of the passive portion of the infalling galaxy population, and so isolates the effect of the final host, quenching times are restricted to around or after the time of the first pericentric passage (marked with a horizontal gray band in the Fig.~\ref{fig_compare}). We also plot for comparison in Fig.~\ref{fig_compare} the result of \citet{2014MNRAS.442.1396W}\footnote{We use the result as reported by \citet{2015arXiv150306803F}, which includes uncertainty estimates.}. Again, the values are not directly comparable with our own, so we attempt to adjust them to match our definitions. We increase their reported timescale by $1.7$~Gyr to account for the difference between the infall time into any more massive host and the most recent infall into a more massive host, again guided by fig.~2 of W13 (the median host mass of the \citealt{2014MNRAS.442.1396W} sample is $10^{13.5}\Msun$), and a further offset of $2.5$~Gyr to account for the travel time between $2.5\,r_{\rm vir}$ and $r_{\rm vir}$\footnote{\citet{2014MNRAS.442.1396W} define infall based on FoF group membership. This definition is unfortunately awkward for comparison; we simply assume that the edge of the FoF group corresponds to $\sim r_{\rm vir}$.}. We omit the other results reported in \citet{2015arXiv150306803F} from our comparison figure as they have no overlap in either $M_{\rm host}$ or $M_*$ with our sample. Recently, a number of authors \citep{2014MNRAS.442.1396W, WetTolWei15, 2015arXiv150306803F, 2016MNRAS.455.2323M} have suggested that galaxies with $M \sim 10^{9} \Msun$ are significantly more resistant to quenching than satellites with both higher and lower masses. Our results for our lowest $M_*$ bin are highly uncertain, but we do seem to find an increase toward $M_* \sim 10^9$, even though we cannot make any strong statements about the timescale at this mass scale. However, the timescales we find at higher masses are clearly lower than the $11.9$~Gyr (value estimated assuming our definitions) found by \citet{2014MNRAS.442.1396W}. \citet{WetTolWei15} infer a timescale of $\sim 8$~Gyr for somewhat less massive $M \sim 10^{8.5} \Msun$ satellites of the Milky Way and M~31, and \citet{2015arXiv150306803F} find substantially lower timescales, again around smaller hosts than those in our sample. This suggests that our results are plausibly consistent with the conclusion that $M_* \sim 10^{9} \Msun$ satellites are most resistant to quenching, but that the host halo mass dependence remains to be better understood. \subsection{Disentangling the physical mechanisms responsible for quenching} It is interesting to consider what the observed timescale of quenching and its dependence on stellar mass reveals regarding the astrophysical mechanisms responsible for quenching star formation. A number of physical processes occur when a galaxy falls into a cluster halo. First, the accretion of dark matter and gas onto the halo is cut off while the satellite halo is still outside the virial radius. Second, ram pressure stripping may strip the hot gas halo as well as the cold gas from the disc. Finally, outflows due to galactic winds may deplete the gas that is available for star formation. It has long been assumed in models of galaxy formation that a galaxy stops accreting gas onto its own halo when it becomes a satellite \citep[e.g.\ ][]{KauWhiGui93, ColAraFre94}, and this is also observed in SPH simulations \citep{KerKatFar09}. The cluster-centric radius at which this cut off occurs is not well known. For example, \citet[][see their fig.~8]{BahMcCBal13} argue that satellite dark matter halos stop growing within $\sim 2 r_{\rm 200c}$ of a cluster, but that their hot gas content is already reduced as far out as $5 r_{\rm 200c}$. \citet{BehWecLu14} have also shown that dark matter accretion ends roughly when the satellite is as far out as $\sim 2 r_{\rm vir}$. So it is likely that the cut-off of accreting gas occurs further out than the fiducial virial radius, and closer to our `backsplash' limit of 2.5 $r_{\rm vir}$. Even if the supply of new gas is cut off, the existing reservoir of cold and hot gas is large enough to sustain star formation in excess of the Hubble time at typical star formation rates: $t \sim (M_{\rm baryon} - M_{*})/{\rm SFR}(M_{*})$. Dividing both numerator and denominator by $M_{*}$ gives $t \sim (f_{\rm baryon}/f_{*}-1)/{\rm SSFR}$, where the fraction of mass in baryons $f_{\rm baryon} \sim 0.15$ and the fraction of total mass in stars $f_{*} \sim 0.01$ for the lowest stellar mass galaxies \citep{HudGilCou15}. Therefore, to explain the short quenching times, additional mechanisms are required to remove or heat the existing gas. As discussed in \S\ref{sec-introduction}, ram pressure stripping of cold gas is clearly seen in galaxy clusters. The key results of this paper are that (i) the quenching occurs approximately at or shortly after pericentre passage, (ii) after the delay $\Delta t$, it is $100$~per~cent effective, (iii) that the time for low-mass galaxies to quench is slightly longer than the time for higher mass galaxies and (iv) the infalling population transitions to become the cluster population relatively quickly (once the delay $\Delta t$ has elapsed), on a timescale $\tau \lesssim 2$~Gyr. Because ram pressure stripping is strongest close to pericentre, the observed timing of the quenching is in broad agreement with the ram pressure stripping model. However, whether the remaining two observations are in accordance with this model is less clear. Ram pressure stripping is expected to be more effective for low mass satellites because the restoring force of the disc is lower, which would argue against the model. However, larger galaxies are more affected by dynamical friction and the ram pressure is very sensitive to speed; it is proportional to the square of the speed through the intra cluster medium. Smaller satellites are likely still proportionally more affected by ram pressure \citep{BahMcC15}, so the trend remains puzzling unless either smaller satellites have some intrinsic property causing them to take longer to cease forming stars or ram pressure is not the dominant trigger of quenching for satellites of all masses in clusters, or some combination of both. \begin{figure} \leavevmode \epsfxsize=\columnwidth \epsfbox{figs/fig10} \caption{Comparison of our quenching timescales as a function of $M_*$ with the simple overconsumption models of \citet{McGBowBal14}, parametrized by the mass-loading factor $\eta$.\label{fig_mcgee}} \end{figure} Recently, \citet{McGBowBal14} have advocated for the combination of a cut-off in gas accretion as a galaxy falls into a cluster, coupled with strong outflows driven by galactic winds (a model which they dub `overconsumption'), to explain the quenching timescales. In their model, the key parameter is the mass-loading of the winds: $\eta = \dot{M}_{\rm out}/{\rm SFR}$, where $\dot{M}_{\rm out}$ represents the rate at which gas is permanently ejected from the satellite's halo. The quenching time is highly sensitive to $\eta$: if $\eta$ is too low then the quenching time is longer than the Hubble time, too high and the quenching time rapidly approaches zero. \citet{McGBowBal14} and \citet{BalMcGMok15} argue that $\eta \sim 1.5$, independent of mass. We have used eq.~7 of \citet{McGBowBal14} to fit their model to our quenching times, assuming no stripping. We have adopted the stellar-to-halo mass relation from weak lensing \citep{HudGilCou15}, and the low redshift SFR of \citet{2007ApJS..173..267S}. Finally, we have also assumed that the `clock' for (over)consumption starts ticking when the satellites crosses $2.5 r_{\rm vir}$. The predictions for contours of constant $\eta$ are shown in Fig.~\ref{fig_mcgee} . The data are fit with a slow varying $\eta$ that ranges from $2.0$ at high mass to $4.0$ at low mass. A slowly-varying $\eta$ model is in conflict, however, with other results on the mass-loading of outflows. In particular, one would expect galaxies with shallower potential wells to have more efficient outflows. This is found in numerical simulations: \citet{MurKerFau15} find that the mass-loading factor scales as $\eta \propto v_{\rm circ}^{-1}$, and \citet{2016arXiv160408244K} finds a constant $\eta\sim 8$ for low mass systems, decreasing with a power law slope of about $-1.8$ above $M_* = 10^{10}\Msun$. From analytic arguments based on the scaling relations of low mass galaxies and the baryonic TF relation, \citet{2012MNRAS.424.3123D} finds that $\eta \sim v_{\rm circ}^{-2}$. These three models would predict higher values of $\eta$ at low stellar mass, and hence short quenching times for low mass galaxies for which winds efficiently remove the gas. The above derived $\eta$ is an upper limit on the true $\eta$, because ram pressure stripping of the hot gas reservoir will substantially reduce the amount of gas potentially available for star formation \citep{BahMcC15}. Furthermore, the `effective' $\eta$ is likely to be higher for infalling satellites than similar counterparts in the field, because weak outflows that in field galaxies would return as a galactic fountain are instead stripped by ram pressure \citep{BahMcC15}. Overall, the timing of quenching near pericentre suggests that ram pressure stripping plays a role. However, the fact that low-mass galaxies have longer quenching time delays than high mass galaxies is difficult to understand, because whatever the mechanism of gas removal, whether ram pressure stripping or galactic winds, it should be more effective in low-mass galaxies with shallower potential wells. We have compared subhalo orbit libraries in projected phase space to star formation rates of SDSS galaxies. This method isolates the environmental effects of the most recent host; in this paper, this is a cluster of mass $10^{13}-10^{15}\Msun$. The key results of this paper are as follows: \begin{enumerate} \item Quenching occurs after a delay time $\Delta t$, measured from first crossing of $2.5r_{\rm vir}$. This delay time is typically $3.5$--$5$~Gyr, with higher mass galaxies quenching slightly earlier. In most cases, this corresponds to times near or shortly after first pericentric approach. All galaxies are quenched on first infall, and before apocentre. \item The delay time does not depend (or depends very weakly) on the host halo mass, over the relatively narrow range probed by our sample. \item Once quenching begins, the timescale $\tau$ for the galaxy population to transition from resembling the galaxies outside the cluster (described by $f_{\rm passive, out}$) to those processed by the cluster (described by $f_{\rm passive, in}$) is fairly short, $\lesssim 2$~Gyr, and usually consistent with $0$~Gyr. Note that this timescale is distinct from the timescale for individual galaxies to transition from an active to a passive state, i.e. the timescale for `crossing the green valley'. \item After the delay has elapsed, the quenching is $100$~per~cent effective, i.e.~all active galaxies that fell in longer ago than $\approx\Delta t + 2\tau$ are passive. The observed fraction of star-forming galaxies in rich clusters is therefore due to a combination of interlopers and galaxies that are falling in for the first time. \end{enumerate} These results appear to be in reasonable agreement with some previous work \citep[W13,][]{2014MNRAS.442.1396W}, after correction the fact that delay times in these works are measured at first accretion and at a different radius. In this paper, we have shown how an orbit library can be used to deproject infalling, backsplash and virialized populations. Here we have compared our projected models with SFR data in PPS, but only for a very simple parameterization of the SFR distribution; this could be extended to leverage the additional information contained in the full SFR distributions. Furthermore, there is no reason to limit the comparison to only SFR, particularly since it is well know that morphology is also correlated with environment. Tidal or harassment effects may also affect the \emph{structures} of discs, possibly stripping them (reduction in stellar mass and radius) or puffing them up so that they are identified morphologically as bulges. In addition, here we have limited ourselves to the correlation between time since infall and PPS position, however much more information is contained in orbit libraries. As redshift surveys continue to improve and grow, we expect that increasingly subtle effects can be teased out of the data.

1607.07934

1607

1607.04106_arXiv.txt

In radio interferometry, observed visibilities are intrinsically sampled at some interval in time and frequency. Modern interferometers are capable of producing data at very high time and frequency resolution; practical limits on storage and computation costs require that some form of data compression be imposed. The traditional form of compression is a simple averaging of the visibilities over coarser time and frequency bins. This has an undesired side effect: the resulting averaged visibilities ``decorrelate'', and do so differently depending on the baseline length and averaging interval. This translates into a non-trivial signature in the image domain known as ``smearing'', which manifests itself as an attenuation in amplitude towards off-centre sources. With the increasing fields of view and/or longer baselines employed in modern and future instruments, the trade-off between data rate and smearing becomes increasingly unfavourable. In this work we investigate alternative approaches to low-loss data compression. We show that averaging of the visibility data can be treated as a form of convolution by a boxcar-like window function, and that by employing alternative baseline-dependent window functions a more optimal interferometer smearing response may be induced. In particular, we show improved amplitude response over a chosen field of interest, and better attenuation of sources outside the field of interest. The main cost of this technique is a reduction in nominal sensitivity; we investigate the smearing vs. sensitivity trade-off, and show that in certain regimes a favourable compromise can be achieved. We show the application of this technique to simulated data from the \ATM{Karl G. Jansky Very Large Array (\JVLA) and the European Very-long-baseline interferometry Network (EVN)}.

\newcommand{\VV}{\mathcal{V}} \newcommand{\PP}{\mathcal{P}} \newcommand{\VVM}{\textcolor{black}{\widehat{\mathcal{V}}}}% \newcommand{\WW}{\mathcal{W}} \newcommand{\II}{\mathcal{I}} \newcommand{\IID}{\mathcal{I}^\mathrm{D}} \newcommand{\IIDI}{\mathcal{I}^\mathrm{DI}} \newcommand{\EE}{\mathcal{E}} \newcommand{\FF}{\mathcal{F}} \newcommand{\HH}{\mathcal{H}} \newcommand{\TT}{\mathcal{T}} \newcommand{\NN}{\mathcal{N}} \newcommand{\uu}{\bmath{u}} \newcommand{\Btf}{\mathsf{B}^{[\Delta t\Delta\nu]}} \newcommand{\Babtf}{\mathsf{B}^{[\alpha\Delta t,\beta\Delta\nu]}} \newcommand{\Bab}{\mathsf{B}^{[\alpha\beta]}} \newcommand{\Buv}{\mathsf{B}^{[uv]}} \newcommand{\Bij}{\mathsf{B}} \newcommand{\Ptf}{\Pi^{[t\nu]}} \newcommand{\Puv}{\Pi^{[uv]}} \newcommand{\Vm}{\textcolor{black}{\widehat{V}}}% \newcommand{\Vs}{V^\mathrm{S}} The following formalism deals with visibilities both as functions (i.e. entire distributions on the $uv$-plane), and single visibilities (i.e. values of those functions at a specific point). To avoid confusion between functions in functional notation and their values, we will use $\VV$ or $\VV(u,v)$ to denote functions, and $V$ to denote individual visibilities. Likewise, $\II(l,m)$ denotes a function on the $lm$-plane i.e. an image. The symbol $\delta$ always denotes the Kronecker delta-function. Depending on whether we want to consider polarisation or not, $\VV$ can be taken to represent either scalar (complex) visibilities, or $2\times2$ complex visibility matrices as per the radio interferometer measurement equation (RIME) formalism \citep{smirnov2011revisiting}. Likewise, $\II$ can be treated as a scalar (total intensity) image, or a $2\times2$ brightness matrix distribution. The derivations below are valid in either case. We shall use the symbols $\mathbf{u}=(u,v)$ or $\mathbf{u}=(u,v,w)$ to represent baseline coordinates in units of wavelength. \subsection{Visibility and relation with the sky} \label{sec:visSky} An interferometer array measures the quantity $\VV(u,v,w)$, known as the visibility function. Here, the coordinates $u,v$ and $w$ are vector components in units of wavelength, describing the distance between antennas $p$ and $q$, called the \emph{baseline}. The $w$ axis is oriented towards the \emph{phase centre} of the observed field, while $u$ points east and $v$ north. Given a sky distribution $\II_0(l,m)$, where $l,m$ are the direction cosines, the nominal observed visibility is given by the van Cittert-Zernike theorem \citep{thompson1999fundamentals,thompson2001fundamentals} as \begin{equation} \VV^\mathrm{nom}(u,v) =\iint\limits_{lm} \frac{\II_0(l,m)}{\sqrt{1-l^2 - m^2}}\,\ee^{-2\pi\ii\phi (u,v,w)}\Rd l\Rd m, \label{eq:visSky:nom} \end{equation} where $\phi(u,v,w)=ul+vm+w(n-1)$, and $n=\sqrt{1-l^2 - m^2}$ (the $n-1$ term comes about when fringe stopping is in effect, i.e. when the correlator introduces a compensating delay to ensure $\phi=0$ at the centre of the field, otherwise the term is simply $n$). Given a pair of antennas $p$ and $q$ forming a baseline $\bmath{u}_{pq}=(u_{pq},v_{pq},w_{pq})$, and taking into account the \emph{primary beam} patterns $\EE_p(l,m)$ and $\EE_q(l,m)$ that define the directional sensitivity of the antennas, this becomes \begin{equation} \VV_{pq}(u,v)=\iint\limits_{lm} \frac{\EE_p \II_0 \EE_q^H}{\sqrt{1-l^2 - m^2}}\,\ee^{-2\pi\ii\phi (u,v,w)}\Rd l\Rd m, \label{eq:visSky} \end{equation} where $^H$ represents the conjugate transpose. The first term being integrated is the \emph{apparent sky seen by baseline} $pq$, \begin{equation} \II_{pq} = \frac{\EE_p \II_0 \EE_q^H}{\sqrt{1-l^2 - m^2}}, \end{equation} which in general can be variable in time and frequency. For simplicity, let us assume that both the sky and the primary beam are constant (invariant in time and frequency), and that the primary beam is the same for all stations. All baselines will then see the same apparent sky throughout the measurement process. Let us designate this by $\II$. Assuming a small FoI ($n\to 1$) and/or a co-planar array ($w=0$), the above equation becomes a simple 2D Fourier transform : \begin{equation} \VV(u,v)=\iint\limits_{lm} \II\,\ee^{-2\pi\ii(ul+vm)}\Rd l \Rd m, \label{eq:visSky:2D} \end{equation} or in functional form, \begin{equation} \VV=\FF\{\II\},~~\II=\FF^{-1}\{\VV\}. \end{equation} \textcolor{black}{we will} refer to $\VV$ as the \emph{ideal} visibility distribution (as opposed to the \emph{measured} distribution, which is corrupted by averaging in the correlator, as \textcolor{black}{we will} explore below). Note that the effect of the primary beam can alternatively be expressed in terms of a convolution with its Fourier transform, the \emph{aperture illumination function} $\mathcal{A}_p(u,v)$. In functional form: \begin{equation} \VV_{pq} = \mathcal{A}_p \circ \VV^\mathrm{nom}_{pq} \circ \mathcal{A}_q^H.\label{eq:visSky:conv} \end{equation} \subsection{Imaging, averaging and convolution} \label{sec:AvgCon} Earth rotation causes the baseline to rotate in time, the baseline in units of wavelength can be treated as a function of frequency and time (from this point on-wards we shall assume that the sky is constant across the range of frequencies being observed): \begin{equation} \label{eq:uvtf} \bmath{u}_{pq}(t,\nu) = \bmath{u}^\mathrm{}_{pq}(t)\nu/c. \end{equation} This, in turn, allows us to rewrite the visibility in eq.~(\ref{eq:visSky:2D}) as a per-baseline function of $t,\nu$: \begin{equation} V_{pq}(t,\nu)=\iint\limits_{lm} \II\,\ee^{-2\pi\ii(u_{pq}(t,\nu)l+v_{pq}(t,\nu)m)}\Rd l \Rd m. \label{eq:visSky:2Dtf} \end{equation} Synthesis imaging recovers the so-called ``dirty image'': the inverse Fourier transform of the measured visibility distribution $\VVM$ sampled by a number of baselines $pq$ at discrete time/frequency points. Inverting the Fourier transform produces the dirty image: \begin{equation} \label{eq:imaging} \IID = \FF^{-1}\{ \WW\cdot\VVM \} , \end{equation} where $\WW$ is the (weighted) sampling function -- a ``bed-of-nails'' function that is non-zero at points where we are sampling a visibility, and zero elsewhere. If $\VVM=\VV$, then this can also be expressed as a convolution of the apparent sky by the \emph{point spread function} $\PP$: \begin{equation} \IID = \PP\circ\II,~~\PP=\FF^{-1}\{\WW\}. \end{equation} Designating each baseline as $pq$, and each time/frequency point as $t_k,\nu_l$, we can represent $\WW$ by a sum of ``single-nail'' functions $\WW_{pqkl}$: \begin{equation} \WW = \sum_{pqkl} \WW_{pqkl} = \sum_{pqkl} W_{pqkl} \delta_{pqkl}, \end{equation} where $\delta_{pqkl}$ is a delta-function shifted to the $uv$-point being sampled: \begin{equation} \delta_{pqkl}(\bmath{u}) = \delta(\bmath{u}-\bmath{u}_{pq}(t_k,\nu_l)) \end{equation} and $W_{pqkl}$ is the associated weight. The Fourier transform being linear, we can rewrite eq.~(\ref{eq:imaging}) as \begin{equation} \label{eq:imaging2} \IID = \sum_{pqkl} W_{pqkl} \mathcal{F}^H\{ \VVM_{pqkl} \}, \end{equation} where \begin{equation} \label{eq:imaging2a} \VVM_{pqkl} = \delta_{pqkl} \Vm_{pq}(t_k,\nu_l) \end{equation} i.e. the visibility distribution corresponding to the single visibility sample $pqkl$. We can further rewrite eq.~(\ref{eq:imaging}) again as \begin{equation} \label{eq:imaging3} \IID = \sum_{pqkl} W_{pqkl} \FF^{-1}\{ \VVM_{pqkl} \}, \end{equation} which shows that the dirty image $\IID$ can be seen as a weighted sum of images corresponding to the individual visibility samples $pqkl$ (each such image essentially being a single fringe pattern). In the ideal case, we would be measuring instantaneous visibility samples, and (assuming no other instrumental corruptions), we would have $\VVM\equiv\VV$, with \begin{equation} \Vm_{pq}(t_k,\nu_l) = \VV(\bmath{u}_{pq}(t_k,\nu_l)), \end{equation} and consequently, \begin{equation} \label{eq:inst-sampled-visibility} \VVM_{pqkl} = \delta_{pqkl} \VV, \end{equation} resulting in what \textcolor{black}{we will} call the \emph{ideal} dirty image $\IIDI$: \begin{equation} \label{eq:imaging:DI} \IIDI = \sum_{pqkl} W_{pqkl} \PP_{pqkl} \circ \II,~~\PP_{pqkl}=\FF^{-1}\{ \delta_{pqkl} \} \end{equation} That is, in the ideal case, each term in the weighted sum is equal to the apparent sky $\II$ convolved with a PSF representing a single visibility sample, $\PP_{pqkl}$. However, an actual interferometer is necessarily non-ideal, in that it can only measure the average visibility over a some time-frequency bin given by the \emph{time} and \emph{frequency sampling intervals} $\Delta t,\Delta \nu$, which \textcolor{black}{we will} call the \emph{sampling bin} \begin{equation} \Btf_{kl} = \bigg [ t_k-\frac{\Delta t}{2},t_k+\frac{\Delta t}{2} \bigg ] \times \bigg [ \nu_l-\frac{\Delta\nu}{2},\nu_l+\frac{\Delta\nu}{2} \bigg ], \label{eq:chap3resamplingbin} \end{equation} This measurement can be represented by an integration: \begin{equation} \Vm_{pqkl} = \frac{1}{\Delta t \Delta \nu} \iint\limits_{\Btf_{kl}} \VV(\bmath{u}_{pq}(t,\nu))\Rd\nu \Rd t. \label{eq2:conti} \end{equation} Inverting the relation of eq.~(\ref{eq:uvtf}), we can change variables to express this as an integration over the corresponding bin $\Buv_{pqkl}$ in $uv$-space: \begin{equation} \Vm_{pqkl} = \frac{1}{\Delta t \Delta \nu} \iint\limits_{\Buv_{pqkl}} \VV_{pq}(u,v)\bigg| \frac{\partial(t,\nu)}{\partial(u,v)}\bigg| \Rd u \Rd v, \label{eq2:conti:uv} \end{equation} where $\Buv_{pqkl}$ is the corresponding bin in $uv$-space. Note that the sampling bins in $t\nu$-space are perfectly rectangular \ATM{(Fig.~\ref{fig:uvcov}, right)} and do not depend on baseline (assuming baseline-independent averaging), while the sampling bins in $uv$-space are elliptical arcs, and do depend on baseline (hence the extra $pq$ index). Assuming a bin small enough that the fringe rate $\partial\bmath{u}/\partial t$ is approximately constant over the bin, we then have \begin{equation} \Vm_{pqkl} \sim \iint\limits_{\Buv_{pqkl}} \VV (\bmath{u}) \Rd\bmath{u}, \label{eq2:conti:uv1} \end{equation} \begin{figure*} \includegraphics[width=0.75\columnwidth]{./Figures/uv_domain_crop.pdf} \includegraphics[width=0.75\columnwidth]{./Figures/tv_domain_crop.pdf} \caption{Schematic of $uv$-coverage for regularly spaced time-frequency samples \ATM{(left: $uv$-space, right: $t\nu$-space).} Baselines with a longer East-West component sweep out longer tracks between successive integration's.}\label{fig:uvcov} \end{figure*} Now, let us introduce a \emph{normalised boxcar window function}, $\Ptf$ \begin{equation} \Ptf(t,\nu) = \bigg \{ \begin{array}{cl} \frac{1}{\Delta t\Delta\nu}, & |t|\leq\Delta t/2,~~|\nu|\leq\Delta\nu/2 \\ 0, & \mathrm{otherwise}, \end{array} \end{equation} using which we may re-write eq.~(\ref{eq2:conti}) as \begin{equation} \Vm_{pqkl} = \iint\limits_{\infty} \VV_{pq}(t,\nu) \Ptf(t-t_k,\nu-\nu_l) \Rd t \Rd \nu, \end{equation} which can also be expressed as a convolution: \begin{equation} \Vm_{pqkl} = [ \VV_{pq} \circ \Ptf ](t_k,\nu_l), \end{equation} Likewise, eq.~(\ref{eq2:conti:uv}) can also be rewritten as a convolution in $uv$-space: \begin{equation} \Vm_{pqkl} = [ \VV_{pq} \circ \Puv_{pqkl} ](\bmath{u}_{pq}(t_k,\nu_l)), \label{eq:avscon} \end{equation} where $\Puv_{pqkl}$ is a boxcar-like window function that corresponds to bin $\Buv_{pqkl}$ in $uv$-space (and also includes the determinant term of eq.~\ref{eq2:conti:uv}). This makes it explicit that each averaged visibility is drawn from a convolution of the underlying visibilities with a boxcar-like window function. Note what eq. (\ref{eq:avscon}) does and does not say. It does say that each individual averaged visibility corresponds to convolving the true visibilities by some window function. However, this window function is different for each baseline $pq$ and time/frequency sample $t_k,\nu_l$ (which is emphasised by the subscripts to $\Puv$ in the equations above). Averaging is thus not a ``true'' convolution, since the convolution kernel changes at every point in the $uv$-plane. \textcolor{black}{We will} call this process a \emph{pseudo-convolution}, and the kernel being convolved with ($\Puv_{pqkl}$) an example of a \emph{baseline-dependent window function} (BDWF). In subsequent sections we will explore alternative BDWFs. In actual fact, a correlator (or an averaging operation in post-processing) deals with averages of discrete and noisy samples, rather than a continuous integration. Ignoring the complexities of correlator implementation, let us cast this process in terms of a simple averaging operation. That is, assume we have a set of \emph{hi-res} or \emph{sampled visibilities} on a high-resolution time/frequency grid $t_i,\nu_j$: \begin{equation} \label{eq:sampling} \Vs_{pqij} = \VV_{pq}(t_i,\nu_j) + \NN[\sigma^\mathrm{(s)}_{pqij}], \end{equation} where $\VV_{pq}$ is given by eq.~(\ref{eq:visSky:2Dtf}), and $\NN$ represents the visibility noise term, which is a complex scalar or complex $2\times2$ matrix with the real and imaginary parts being independently drawn from a zero-mean normal distribution with the indicated r.m.s. \citep{wrobel1999sensitivity}. The noise term is not correlated across samples. The \emph{lo-res} or \emph{averaged} or \emph{resampled} visibilities are then a discrete sum: \begin{equation} \label{eq:discrete:tf0} \Vm_{pqkl} = \frac{1}{n} \sum_{ij\in\Bij_{kl}} \Vs_{pqij}, \end{equation} where $\Bij_{kl}$ is the set of sample indices $ij$ corresponding to the \emph{resampling bin}, i.e. \begin{equation} \Bij_{kl} = \big \{ ij:~t_i\nu_j \in \Btf_{kl} \big \}, \end{equation} and $n = n_t\times n_\nu$ is the number of samples in the bin. Using the BDWF definitions above, this becomes a conventional discrete convolution (assuming a regular $t\nu$ grid): \begin{equation} \label{eq:discrete:tf} \Vm_{pqkl} = \sum_{i,j=-\infty}^{\infty} \Vs_{pqij} \Ptf(t_i-t_k,\nu_j-\nu_l). \end{equation} In $uv$-space, this becomes a discrete convolution on an irregular grid (the $\bmath{u}_{ij}$ grid being schematically illustrated by \ATM{Fig.~\ref{fig:uvcov}, left)}: \begin{equation} \label{eq:discrete:uv} \Vm_{pqkl} = \sum_{i,j=-\infty}^{\infty} \Vs_{pqij} \Puv_{pqkl}(\bmath{u}_{ij}-\bmath{u}_{kl}), \end{equation} \subsection{Effect of averaging on the image} \label{sec:effectbw} \begin{figure*} \includegraphics[width=.4\textwidth]{./Figures/idealIPRgrey.pdf}% \includegraphics[width=.4\textwidth]{./Figures/idealsincgrey.pdf}\\ \caption{Left: boxcar response. In the $uv$-plane, this represents the window function corresponding to normal averaging of visibilities. In the image plane, this represents the ideal image-plane response function. Right: \Sincc~response. In the image plane, this represents the window response function corresponding to a boxcar window function in the $uv$-plane. In the $uv$-plane, this represents the ideal window function.} \label{fig:idealwindowfunction} \end{figure*} In the limit of $\Delta t,\Delta \nu \rightarrow 0$, averaging becomes equivalent to sampling. An interferometer must, intrinsically, employ a finitely small averaging interval. The Fourier phase component $2\pi\phi(u,v,w)$ is a function of frequency and time, with increasing variation over the averaging interval for sources far from the phase centre. The average of a complex quantity with a varying phase then effectively ``washes out'' amplitude, the effect being especially severe for wide FoIs \citep[for an extensive discussion, see][]{bregman2012system}. In the $uv$-plane, this effect is often referred to as \emph{time} and \emph{bandwidth decorrelation}, and \emph{smearing} in the image plane. The discussion above provides an alternative way to look at decorrelation/smearing. With averaging in effect, the relationship between the measured and the ideal visibility changes to (contrast this to eq.~\ref{eq:inst-sampled-visibility}): \begin{equation} \VVM_{pqkl} = \delta_{pqkl} ( \VV\circ\Puv_{pqkl} ), \end{equation} Combining this with eq.~\ref{eq:imaging3}, and using the Fourier convolution theorem, we can see that the dirty image is formed as\textcolor{black}{:} \begin{equation} \IID = \sum_{pqkl} W_{pqkl} \PP_{pqkl} \circ (\II\cdot\TT_{pqkl}), \end{equation} with the apparent sky $\II$ now tapered by the baseline-dependent \emph{window response function} $\TT_{pqkl}$, the latter being the inverse Fourier transform of the BDWF: \begin{alignat}{2} \TT_{pqkl} &= \FF^{-1}\{ \Puv_{pqkl} \}. \end{alignat} In other words, the dirty image yielded by averaged visibilities (compare this to the ideal dirty image given by eq.~\ref{eq:imaging:DI}) is a weighted average of per-visibility dirty images corresponding to a per-visibility tapered sky. The Fourier transform of a boxcar-like function is a \Sincc-like function, schematically illustrated in 1-D by Fig.~\ref{fig:idealwindowfunction} (right). Time and bandwidth smearing represents the average effect of all these individual tapers. Shorter baselines correspond to smaller boxcars and wider tapers, longer baselines to larger boxcars and narrower tapers, and are thus more prone to smearing. \begin{figure*} \includegraphics[width=\columnwidth]{./Figures/effect_time_averaging.pdf}% \includegraphics[width=\columnwidth]{./Figures/effect_bandwidth_averaging.pdf} \caption{Effects of time and frequency averaging: the apparent intensity of a 1 Jy source, as seen by \JVLA-C at 1.4 GHz, as a function of distance from phase centre. (Left) Frequency interval fixed at 125 kHz, time interval varies; (right) time interval fixed at 1 s, frequency interval varies. }\label{fig:smear} \end{figure*} Fig.~\ref{fig:smear} (produced by simulating a series of high time-frequency resolution observation using MeqTrees~\citep{noordam2010meqtrees}, and applying averaging) shows the attenuation of a 1 Jy source as a function of distance from phase centre, for a set of different time and frequency intervals. The simulations correspond to \JVLA~ in the C configuration, with an observing frequency of 1.4 GHz. At this frequency, the first null of the primary beam is at $r\approx36'$, and the half-power point is at $\sim16'$, thus we can consider the ``conventional'' FoI (i.e. the half-power beam width, or HPBW) to be about $0.5^\circ$ across. Note that the sensitivity of the upgraded \JVLA, as well as improvements in calibration techniques \citep{Perley-3C147}, allow imaging to be done in the first primary beam sidelobe as well (and in fact it may be necessary for deep pointing, if only to deconvolve and subtract sidelobe sources), so we could also consider an ``extended'' FoI out to the second null of the primary beam at $r\approx1.25^\circ$. Whatever definition of the FoI we adopt, Fig.~\ref{fig:smear} shows that to keep amplitude losses across the FoI to within some acceptable threshold, say 1\%, the averaging interval cannot exceed some critical size, say 10 s and 1 MHz. Conversely, if we were to adopt an aggressive averaging strategy for the purposes of data compression, say 50 s and 5 MHz, the curves indicate that we would suffer substantial amplitude loss towards the edge of the FoI. Finally, note that the curves corresponding to acceptably low values of smearing across the FoI (i.e. up to 25 s and up to 1.25 MHz) have a very gentle slope, with very little suppression of sources \emph{outside} the FoI. \subsection{The case for alternative BDWFs} The window response or image plane response (IPR) function induced by normal averaging (Fig.~\ref{fig:idealwindowfunction} (right)) is far from ideal: it either suppresses too much within the FoI, or provides limited benefit to suppressing outside the FoI sources, or both. The optimal image plane response would be a disk-like function, unity within the region bounded by a circle and zero outside this region. In 1-D, it is a boxcar-like, as in Fig.~\ref{fig:idealwindowfunction} (left). The BDWF that would produce such a response is \Sincc-like, as in 1-D presented in Fig.~\ref{fig:idealwindowfunction} (right). The problem with a \Sincc~is that it has infinite support; applying it over finite-sized bins necessarily means a \emph{truncated} BDWF that results in a sub-optimal taper. The problem of optimal filtering has been well studied in signal processing (usually assuming a true convolution rather than the pseudo-convolution we deal with here), and we shall apply these lessons below. The derivations above make it clear that using a different BDWF in place of the conventional boxcar-like $\Puv$ could in principle yield a more optimal tapering response. The obvious \textcolor{black}{disadvantage} is a loss in sensitivity. Each visibility sample is subject to an independent Gaussian noise term in the real and imaginary part; the noise of the average of a set of samples is minimised when the average is naturally weighted (or unweighted, if the noise is constant across visibilities). Thus, any deviation from a boxcar F must necessarily increase the noise in the visibilities. Below we will study this effect both theoretically and via simulations, to establish whether this trade-off is sensible, and under which conditions.

The goal of this work was to demonstrate the application of baseline-dependent window functions to radio interferometry. We have demonstrated that BDWFs offer a number of interesting advantages over conventional averaging. The first of these is data compression -- i.e. visibilities can be sampled at a lower rate, while retaining a large FoI. Compression by a factor of 16 with relatively little loss of sensitivity has been demonstrated. The huge data rates from upcoming instruments such as ASKAP, MeerKAT and the future SKA1 mean that raw visibilities may need to be discarded after calibration (unlike older instruments, where raw visibilities have typically been archived). This represents something of a risk to the science, as it precludes future improvements in calibration techniques from being later applied to the data. With BDWFs, at least a highly-compressed version of the visibilities may be retained. The second potential benefit of BDWFs is the increased suppression of unwanted signal from out-of-FoI sources. This reduces both the overall level of far sidelobe confusion noise, and lessens the impact of A-team sources in sidelobes. Thirdly, BDWFs can have an interesting impact in the VLBI case, as they allow the full primary beam FoI to be imaged using a single VLBI dataset. This opens the door to wide-field VLBI, which has previously been impractical. BDWFs have a number of potential downsides. The first one is a potential loss in sensitivity. Our simulations show that this can be kept within reasonable limits, especially if overlapping BDWFs are employed, and can be traded off with compression rate. Note that from DSP theory, we know that a properly matched filter can actually increase the SNR: the equivalent result with BDWFs is that off-axis sources are attenuated less, so the effective SNR off-axis can increase despite the loss in absolute sensitivity. The second downside of BDWFs is an increase in computational complexity. Whether implemented in a correlator or in post-processing, BDWFs (and especially overlapping BDWFs) require substantially more operations than simple averaging. There may be other limits to the practical applicability of BDWFs. They are far less efficient if high spectral resolution is required, so their use may be limited to continuum observations. Furthermore, averaging over longer intervals requires accurate phase calibration, so high compression rates may only be achievable post-calibration. An interesting avenue of future research is combining BDWFs with baseline-dependent averaging. As we saw above, the ability of BDWFs to shape a FoI is somewhat limited by the fact that shorter baselines sweep out smaller bins in $uv$-space, with window functions over them becoming boxcar-like. If baseline-dependent averaging is employed, shorter baselines are averaged over larger $uv$-bins, thus increasing the effect of BDWFs. Finally, we should note that the use of BDWFs results in a different position-dependent PSF than regular averaging (or to put it another way, the smearing response of BDWFs results in a different smeared PSF shape). Future work will focus on methods of deriving this PSF shape, with a view to incorporating this into current imaging algorithms.

1607.04106

1607

1607.01453_arXiv.txt

All of the 14 subfields of the \kepler\ field have been observed at least once with the Large Sky Area Multi-Object Fiber Spectroscopic Telescope (\lamost, Xinglong Observatory, China) during the 2012-2014 observation seasons. There are 88,628 reduced spectra with \snrg\ (signal-to-noise ratio in $g$ band) $\geq$ 6 after the first round (2012-2014) of observations for the \lamost-\kepler\ project (\project). By adopting the upgraded version of the \lamost\ Stellar Parameter pipeline (\lasp), we have determined the atmospheric parameters (\teff, \logg, and \feh) and heliocentric radial velocity \vrad\ for 51,406 stars with 61,226 spectra. Compared with atmospheric parameters derived from both high-resolution spectroscopy and asteroseismology method for common stars in \citet{Hub2014}, an external calibration of \lasp\ atmospheric parameters was made, leading to the determination of external errors for the giants and dwarfs, respectively. Multiple spectroscopic observations for the same objects of the \project\ were used to estimate the internal uncertainties of the atmospheric parameters as a function of \snrg\ with the unbiased estimation method. The \lasp\ atmospheric parameters were calibrated based on both the external and internal uncertainties for the giants and dwarfs, respectively. A general statistical analysis of the stellar parameters leads to discovery of 106 candidate metal-poor stars, 9 candidate very metal-poor stars, and 18 candidate high-velocity stars. Fitting formulae were obtained segmentally for both the calibrated atmospheric parameters of the \project\ and the KIC parameters with the common stars. The calibrated atmospheric parameters and radial velocities of the \project\ will be useful for studying stars in the \kepler\ field.

The main scientific objective of the NASA (National Aeronautics and Space Administration) space mission \kepler\ are to detect the Earth-size and even larger planets in the habitable zone \citep{Kas1993,Bor2007} of solar-like stars by using the method of photometric transits \citep{Bor2009}, and to determine the properties of the planet host stars by means of asteroseismic methods \citep{Chr2007}. Since the successful launch of \kepler\ on March 7, 2009, the number of stars having photometric time-series with an ultra-high precision of a few micro-magnitudes has increased steadily over a time-span of 4 years. As a large number of uninterrupted time-series has been obtained for pulsating stars of all kinds and flavors, the \kepler\ mission provides an unprecedented opportunity to study stellar oscillations. The \kepler\ Asteroseismic Science Consortium (\kasc, \citealt{Chr2007}), with a broad community participation, was established to select the most promising asteroseismic targets in the \kepler\ field of view (hereafter `\kepler\ field') as targets for \kepler\ and to study their internal structure by means of asteroseismic methods \citep{Gil2010,Cha2010}. However, a reliable asteroseismic modeling requires reliable basic stellar physical parameters such as atmospheric parameters (the effective temperature \teff, the surface gravity \logg, and the metallicity \feh) and the projected rotational velocity (\vsini). Unfortunately, the atmospheric parameters as given in the \kepler\ Input Catalogue (KIC, \citealt{Bro2011}) are not always unsuited for a successful asteroseismic modeling as their errors amount to $\sim$200 K in \teff, and to $\sim$0.5 dex in both \logg\ and \feh. Moreover, KIC atmospheric parameters are missing for a significant fraction of the \kepler\ objects. The shortcomings of the stellar properties in the KIC have been quantified in follow-up studies and are summarized by \citet[hereafter H14]{Hub2014}. During the last decade, a lot of ground-based observations have been gathered for a wide variety of \kepler\ targets to support their space-based observations. An enormous observational effort involving 2-m class telescopes located in 12 countries in the northern hemisphere has been coordinated by \citet{Uyt2010a,Uyt2010b} for the observations of \kasc\ objects, which leads to the characterization of, amongst others, OB-type stars \citep[including candidate $\beta$\,Cephei and slowly pulsating B stars;][]{Cat2010, Leh2011,Tka2013}, AF-type stars \citep[including candidate $\delta$\,Scuti and $\gamma$\,Doradus stars;][]{Cat2011,Tka2012,Tka2013,Nie2015}, solar-like stars \citep{Mol2008,Bru2012,Kar2013}, giants \citep{Bru2011}, and red giants \citep{Thy2012}. Though strong efforts have been made to characterize all types of asteroseismic targets, a significant fraction of the \kasc\ targets remained unobserved, mainly because of the faintness of the targets and the unavailability of a sufficient amount of telescope time. H14 presented an improved catalog for 196,468 stars observed by the NASA \kepler\ mission to support the study of the planet-occurrence rate by consolidating the published values of the atmospheric parameters (\teff, \logg, \feh) that are derived with different observational techniques (mainly photometry, spectroscopy, asteroseismology and exoplanet transits). It is a valuable contribution to the improvement of the stellar properties of \kepler\ targets, but for a considerable fraction of stars, the KIC parameters could not be updated. Moreover, the consistency in the results is lacking as they are based on observations from heterogeneous devices and analysis techniques. The \lamost\ (the Large Sky Area Multi-Object Fiber Spectroscopic Telescope, also called the GuoShouJing Telescope) \citep{Su1998,Zhao2012}, is a special 4-meter reflecting Schmidt telescope located at the Xinglong station of the National Astronomical Observatories of China \citep{Cui2012,Luo2012}. Its focal length is 20 m and the focal plane, with a diameter of 1.75 m corresponding to a circular field of view of 5 degrees on the sky \citep{Wang1996}, is covered with 4000 optical fibres connected to 16 two-arm low-resolution spectrographs with 32 CCD cameras. \lamost\ spectra have a resolution of about 1800 and cover the wavelength range 370-900 nm \citep{Cui2012,Zhao2012}. The combination of a large aperture and a wide field of view covered by 4000 fibers makes \lamost\ the most powerful optical spectroscopic survey instrument in the northern hemisphere at present. We therefore initiated the \lk\ project \citep[\project;][]{Dec2014} to acquire \lamost\ spectra for as many objects in the \kepler\ field as possible and to characterize them in terms of spectral classification (spectral type with any peculiarities), atmospheric parameters (\teff, \logg, \feh), rotation rate (\vsini) and radial velocity (\vrad). It is the only way we can derive these parameters with an accuracy required for a detailed asteroseismic study for the vast majority of the \kepler\ objects in an efficient and homogeneous way. For a detailed description of the \project, we refer interested readers to \citet{Dec2015}. The low-resolution \lamost\ spectra available in the catalogue of the \project\ \citep{Dec2015} have been analysed by three different teams, each with their own independent method. This paper presents the analysis of the released parameters from the `Asian team' (composed by the \lamost\ project's data processing department group and ABR, JNF, XHY group from Beijing Normal University) who determined the stellar atmospheric parameters by using the official \lamost\ Stellar Parameter pipeline \citep[\lasp,][]{Wu2014,Wu2011a,Luo2015}. This paper is organized as follows. In Sections \ref{sect:2} and \ref{sect:3}, brief descriptions of the observations and spectral data are given, respectively. A concise introduction of the \lasp\ stellar parameter calculation is given in Section \ref{sect:4}. In Section \ref{sect:5}, \lasp\ stellar parameters and their errors are calibrated based on the results from both external calibrations and internal uncertainties for the giants and dwarfs, respectively. Then we perform a statistical analysis of the calibrated atmospheric parameters and radial velocities present in the \lasp\ catalogue in Section \ref{sect:6}. It includes identifications of candidates of particular objects. We compare calibrated \lasp\ parameters with the values in KIC for common stars in Section \ref{sect:7}. The paper is ended with conclusions and the prospect of the \project\ in Section \ref{sect:8}.

1607.01453

1607

1607.03984_arXiv.txt

{The influence of the anisotropy in the equilibrium and stability of strange stars is investigated through the numerical solution of the hydrostatic equilibrium equation and the radial oscillation equation, both modified from their original version to include this effect. The strange matter inside the quark stars is described by the MIT bag model equation of state. For the anisotropy two different kinds of local anisotropic $\sigma=p_t-p_r$ are considered, where $p_t$ and $p_r$ are respectively the tangential and the radial pressure: one that is null at the star's surface defined by $p_r(R)=0$, and one that is nonnull at the surface, namely, $\sigma_s=0$ and $\sigma_s\neq0$. In the case $\sigma_s=0$, the maximum mass value and the zero frequency of oscillation are found at the same central energy density, indicating that the maximum mass marks the onset of the instability. For the case $\sigma_s\neq0$, we show that the maximum mass point and the zero frequency of oscillation coincide in the same central energy density value only in a sequence of equilibrium configurations with the same value of $\sigma_s$. Thus, the stability star regions are determined always by the condition $dM/d\rho_c>0$ only when the tangential pressure is maintained fixed at the star surface's $p_t(R)$. These results are also quite important to analyze the stability of other anisotropic compact objects such as neutron stars, boson stars and gravastars.} \begin{document}

In astrophysics, the study of compact stars is vitally important because it provides an excellent laboratory for the study of highly dense matter on extreme conditions. Theoretically, it is largely considered that these objects are composed of a perfect fluid. However, strong theoretical evidences suggest that in a highly dense fluid, as the one contained in a compact star, different physical phenomena could give rise to anisotropy. For example, because of the geometry of the $\pi^{-}$ modes, anisotropic distribution of pressure could be considered in a neutron star to describe a pion condensed phase configuration \cite{sawyer}. Anisotropy could be caused within a neutron star due to the presence of a solid or a superfluid core \cite{ruderman1972,heiselberg2000}. It appears in self-gravitating objects determined by complex scalar fields, i.e., boson stars, \cite{gleiser1988,gleiser1989}. In addition, the anisotropy could be of importance in other context, for instance, to explore the hypothesis that the event GW$150914$ \cite{Gravi_waves} could be produced by the merging of rotating gravastars and not by black holes \cite{chirenti_rezzolla2016}. Independently of the nature of the anisotropy, it might produce considerable changes in the physical properties of compact stars. In the scope of General Relativity we can find a vast number of works studying the influence of the anisotropy on static spherically symmetric objects composed of a perfect fluid (review \cite{mak_harko2003,herrera_santos1997} and references therein). In order to see the influence of the anisotropy on physical properties of spherically symmetric objects made of an isotropic fluid, the anisotropic stress tensor is added to the perfect fluid energy-momentum tensor. The components of the resulting energy-momentum tensor $T^{1}_{1}$ and $T^{2}_{2}=T^{3}_{3}$ are usually renamed as the radial and tangential pressure, $p_r$ and $p_t$ respectively. The difference between the pressures $p_r$ and $p_t$ is what gives rise to the anisotropy of a fluid. The difference between these two pressures is known as anisotropic factor $\sigma=p_t-p_r$. All pressures in question are functions of the radial coordinate. Thus the spherical symmetry is preserved, since we do not have inside the star tangential forces perpendicular to the radial direction. The first study of the influence of anisotropic compact objects was analyzed by Bowers and Liang \cite{bowers_liang1974} in $1974$. In that work, the importance of the anisotropy of incompressible stars (stars with constant energy density, known also as Schwarzschild star) is studied through the generalization of the hydrostatic equilibrium equation, modified from it's original form to include the anisotropy effects. Using the anisotropic factor $\sigma=\alpha\,(p_r+\rho)\,(\rho+3\,p_r)\,e^{\lambda}\,r^n$, $\alpha$ and $n$ being constants (with $n>1$) and $e^{\lambda}$ a metric function, they found that the anisotropy has considerable effects on the maximum mass and on the surface redshift. As argued by the authors, the total radius of an incompressible star will be determined by the condition $p_r=0$, even though the anisotropic factor is nonnull on the surface, $\sigma_s\neq0$. In other words, an equilibrium solution can be found even with $\sigma_s\neq0$. In this work the radial stability is not investigated. Shortly later, the effect of anisotropy on several macroscopic properties of neutron stars was analyzed by Hillebrandt in collaboration with Heintzmann in \cite{hillebrandt1975} and together with Steinmetz in \cite{hillebrandt1976}, considering that the anisotropic factor follows the function $\sigma= \beta\,p_r$, $\beta$ being a constant. Heintzmann and Hillebrandt in \cite{hillebrandt1975} found that for an arbitrarily large anisotropy there is no limiting mass and nor limiting redshift for neutron stars. After generalizing the Chandrasekhar radial pulsation equation for an object composed of an anisotropic fluid, Hillebrandt and Steinmetz in \cite{hillebrandt1976} analyzed the stability of the neutron stars. They show that a criterion in the stability against radial pulsation exists, similar to the one determined for isotropic models. In recently years, the influence of anisotropy both in nonradial oscillations, as in slowly rotating neutron stars, has been investigated considering the anisotropic factors $\sigma = \beta\,p_r(1-e^{ \lambda})$ in \cite{doneva_yazadjiev}, and with the one used in \cite{bowers_liang1974} (with $n=2$), in Ref. \cite{silva_macedo_berti_crispino2015}. Doneva and Yazadjiev in \cite{doneva_yazadjiev} analyze the influence of the anisotropy in the oscillation spectrum of a neutron star in the Cowling approximation. In that work, the authors found that the anisotropy could play an important role in the oscillation spectrum of a neutron star. Silva {\it et al.} in \cite{silva_macedo_berti_crispino2015} investigate the effects of the anisotropy on slowly rotating stars in General Relativity and in scalar-tensor theory. Silva and collaborators found that the anisotropy affects some physical properties of the neutron stars, such as the moment of inertia (in both theories) and the scalarization (a phase transition similar to spontaneous magnetization in ferromagnetic materials). The authors determined that the effects of scalarization grow (shrink) when the anisotropy increases (decreases). In an anisotropic neutron star, unlike an anisotropic incompressible star, both the radial pressure and the energy density are zero on its surface, requiring that the anisotropy, in any of the forms previously given, is null on the star's surface $\sigma_s=0$. From the afore cited works, we see that two possible forms of anisotropic factors can be analyzed, those that do not vanish on the star's surface $\sigma_s\neq0$ and those that do $\sigma_s=0$. In this work, we study for the first time the stability of anisotropic strange stars in these two cases. As is known, if strange matter is the true ground state \cite{witten1984}, the compact objects known as pulsars would be strange stars rather than neutron stars \cite{itoh1970, Negreiro2009, Weber2007, Malheiro2003}. This leads to the need of studying anisotropic effects also in these types of compact objects. Possible sources of anisotropic pressure in quark matter could be color supercoductivity and viscosity, in the same way as superfluid and viscosity are in neutron stars, as has been discussed before. We assume that the radial pressure and fluid energy density are connected by the MIT bag model equation of state. This equation of state was used in previous works, for instance, to study the radial oscillations of strange stars \cite{vath_chanmugam1992,benvenuto_horvath1991,gondek1999} and the stability of thin shell interfaces inside compact stars \cite{pereira_coelho_rueda2014} (review also \cite{pereira_rueda2015}). In turn, the anisotropic equation of state will be described by two possibles cases, one that follows the function $\sigma=\alpha\,(p_r+\rho)\,(\rho+3\,p_r)\,e^{\lambda}\,r^2$ and another one of the form $\sigma = \beta\,p_r(1-e^{ \lambda})$. Note that in a strange star, where the fluid is described by the MIT bag model equation of state, the energy density at the surface of the star is nonnull. This allows the first and second anisotropy factor to be respectively nonzero and zero at the star's surface, i.e., $\sigma_s\neq0$ and $\sigma_s=0$, respectively. The article is organized as follows. In Sec.~\ref{gen_re_eq} we present the stress-energy tensor, both the stellar structure equations and the radial pulsation equations and their boundary conditions. In Sec.~\ref{eos_anis_carga} we make a brief description of the equation of state and of the two anisotropic factors considered. In Sec.~\ref{results} we present the numerical method used to integrate the stellar structure equations and the radial oscillation equations. We also present the scaling solution and results about the influence of the anisotropy in equilibrium and on the stability of strange quark stars for the two cases considered. Finally, in Sec.~\ref{conclusion} we conclude. It is worth mentioning that in this work the units $c=1=G$, where $c$ is the speed of light and $G$ is the gravitational constant, are considered. The metric is of signature $+2$ and Greek indices run from $0$ to $3$.

The radial stability of anisotropic strange quark stars is studied in the present article. This is possible through the numerical solution of the hydrostatic equilibrium equation and the radial pulsation equation, which are modified from their original form to the inclusion of anisotropy. We consider that the matter contained in the strange stars follows the MIT bag model equation of state. On the other hand, two different kinds of anisotropic factor $\sigma$ are considered to describe the anisotropy. One that is nonvanishing on the surface of the star and other one that vanishes on it, namely, $\sigma_s\neq0$ (in case $1$) and $\sigma_s=0$ (in case $2$), respectively. Specifically, the anisotropic factors considered take the form $\sigma=\frac{\kappa}{1.7}(\rho+p_r)(\rho+3\,p)r^2 e^{\lambda}$, for the case $1$, and $\sigma=\kappa p_r (1-e^{-\lambda})$, for the case $2$. In the two cases analyzed, we found that the anisotropy affects physical properties of the stars, such as: the energy density, the radial pressure, the total radius, the total mass, the surface redshift, and the frequency of oscillation of the fundamental mode. In case $1$, when $\sigma_s\neq0$, we found that the maximum mass point and the zero frequency of oscillation are obtained for different central energy densities. From this we understand that the conditions $dM/d\rho_c>0$ and $dM/d\rho_c<0$ are necessary but not sufficient to recognize regions corresponding to stable and unstable stars in a sequence of equilibrium configurations with fixed $\kappa$. In turn, in the case $2$, when $\sigma_s=0$, the maximum mass point and $\omega=0$ are found at the same central energy density. This indicates that the maximum mass point marks the onset of the instability in this case. In other words, when $\sigma_s=0\ (p_t(R)=p_r(R)=0)$, the conditions $dM/d\rho_c>0$ and $dM/d\rho_c<0$ can be used to distinguish between a stable and an unstable region in a sequence of equilibrium configurations with fixed $\kappa$. In case $1$ (when $\sigma_s\neq0$), the maximum mass point and the zero frequency of oscillation are obtained at the same central energy density only in a system of equilibrium configurations with the same $\sigma_s \ (p_t(R)\neq 0\ \textrm{and fixed})$. This indicates that in such sequences with $\sigma_s$ fixed, the conditions $dM/d\rho_c>0$ and $dM/d\rho_c<0$ can be used to determine the regions constituted by stable or unstable against radial perturbations, respectively. In addition, we determine that for a range of some parameters, the anisotropy $\sigma_s$ influences the stability of the star. For a range of $\rho_c$, when the anisotropy at star surface $\sigma_s> 0$, the growth of $\sigma_s$ reduces the stability of the star. Moreover, for a range of masses, while $\sigma_s>0$, the increment of $\sigma_s$ increases the stability of the star. \

1607.03984

1607

1607.04040_arXiv.txt

We investigate a correlation between star-formation rate (SFR) and stellar mass for \ha\ emission line galaxies (HAEs) in one of the richest proto-clusters ever known at $z\sim2.5$, USS~1558-003 proto-cluster. This study is based on a 9.7-hour narrow-band imaging data with MOIRCS on the Subaru telescope. We are able to construct a sample, in combination with additional $H$-band data taken with WFC3 on Hubble Space Telescope (HST), of 100 HAEs reaching the dust-corrected SFRs down to 3 \Msun\ yr$^{-1}$ and the stellar masses down to $10^{8.0}$ \Msun. We find that while the star-forming galaxies with $\ga10^{9.3}$ \Msun\ are located on the universal SFR-mass main sequence irrespective of the environment, less massive star-forming galaxies with $\la10^{9.3}$ \Msun\ show a significant upward scatter from the main sequence in this proto-cluster. This suggests that some less massive galaxies are in a starburst phase, although we do not know yet if this is due to environmental effects.

Since the last decade, the question of a positive correlation between SFR and stellar mass in star-forming galaxies (SFGs), which is called the main sequence (MS) of SFGs, has been one among hot topics in the field of galaxy evolution \citep[e.g.,][]{Noeske2007,Daddi2007,Elbaz2007}. The tight correlation provides us with perspectives of how the SFGs evolve over cosmic time: They spend most of their lifetimes on the sequence and evolve along the MS. However, a small fraction of them shows starburst activities, and they deviate upwards from the MS \citep{Rodighiero2011}. During the course of hierarchical structure formation, galaxy evolution is expected to proceed in different ways in different environments. It is suggested that such environmental effects are more preferentially seen in satellite galaxies rather than in central galaxies at $z<1$ \citep[e.g.,][]{Peng2012,Kovac2014}. Some environment-dependent processes such as galaxy interactions/merging and gas inflows/outflows can alter the star-formation activity in galaxies, either boosting it or truncating it. Understanding the physical mechanisms of these processes is of vital importance to reveal the origin of early-type galaxies and the strong environmental dependence of galaxy properties seen in the present-day Universe. With this motivation, we have been conducting a systematic project called MAHALO-Subaru (MApping H-Alpha and Lines of Oxygen with Subaru; \citet{Kodama2013}) and mapping star-formation activities over a wide range of environments and across cosmic times, in particular at $1.5 \lesssim z \lesssim 2.5$, where clusters of galaxies are just assembling and galaxies are forming vigorously therein. The project has shown that integrated SFR per dynamical mass in cluster core increases dramatically with redshift up to $z\sim2.5$ \citep{Shimakawa2014}. However, \citet{Koyama2013b} show that the location of the MS of SFGs in a proto-cluster (PKS1138--262) at $z\sim2$ is not different from that in the general field at similar redshifts, although the distribution of galaxies along the MS is skewed to higher SFRs and stellar masses in high density regions probably due to biased, more advanced galaxy formation there. Since our analyses have been limited to relatively massive galaxies ($\ga10^{9.5}$ \Msun) so far, we want to extend such study to an even less massive regime. For this purpose, we target a proto-cluster around the USS~1558-003 radio galaxy at $z=2.53$. \citet[][hereafter \hayashi12]{Hayashi2012} have already reported a narrow-band \ha\ emission line survey as a part of MAHALO-Subaru. This previous observation has identified as many as 68 HAEs associated with the proto-cluster within a 27 arcmin$^2$ field of view (FoV). It shows a linear filamentary structure which hosts three dense groups of HAEs. The richness and high density make it a unique proto-cluster target at $z>2$ for us to investigate the early environmental effects in the galaxy-formation phase. To access less massive galaxies in the cluster, we have conducted very deep follow-up observations: One is three times deeper narrow-band \ha\ imaging with Subaru Telescope. Another is deep HST/WFC3 imaging at near-infrared. In this {\it Letter}, based on these new unique imaging data-sets, we report the first intriguing discovery of the nature of the less massive SFGs ($\la10^{9.5}$ \Msun) in this proto-cluster USS~1558-003 at $z=2.53$, which has become accessible only with these deep observations. Magnitudes are presented in the AB system \citep{Oke1983}, cosmological parameters of ${\rm H}_0=70$ km s$^{-1}$ Mpc$^{-1}$, $\Omega_{m}=0.3$ and $\Omega_{\Lambda}=0.7$, and \citet{Chabrier2003} initial mass function, are adopted throughout the letter.

\label{sec:discussion} Until now, there have been a number of previous studies that investigated the MS of SFGs at $z\ga2$ \citep[e.g.,][]{Dunne2009,Karim2011,Reddy2012}. Figure \ref{fig:MS} also compares the MSs derived from the previous studies. Our HAEs with stellar masses of $\ga10^{10}$ \Msun\ seem to be located right on the previously measured MSs in the literature \citep{Whitaker2014,Shivaei2015,Speagle2014}. Since these previous studies mainly look at galaxies in the general fields, the agreement suggests that SFGs in the proto-cluster at $z\sim2.5$ share the same MS as field SFGs at similar redshifts, consistent with the previous studies \citep{Koyama2013b, Cooke2014}. On the other hand, at lower stellar masses ($<10^{9.3}$ \Msun) in the proto-cluster, there are several galaxies that are significantly up-scattered above the MSs. If such small mass galaxies all follow the same extrapolated MS, they would be all located below our detection limit and we would not see any of them, contrary to what we actually see. Although we cannot discuss the exact locations of the MS at $<10^{9.3}$ \Msun\ due to incompleteness, we argue that there are at least some HAEs that are significantly deviated upward from the MS. Those more than 10 up-scattered HAEs at $<10^{9.3}$ \Msun\ have exceptionally large specific SFRs (sSFR=SFR/\Mstar) above 10$^{-8}$ yr$^{-1}$ as shown by the light solid line in Figure~\ref{fig:MS}. This indicates that their inferred ages (timescales of star formation) are smaller than 10$^8$ years, and they are young starbursting galaxies just being formed. Note that this result is not affected by dust corrections, because the amount of dust correction is progressively lower for less massive galaxies \citep{GarnBest2010,Koyama2015}. In fact, the inferred A(\ha) for almost all of the galaxies with $<10^{9}$ \Msun\ are smaller than 0.2 mag. Even if we use the rest-frame UV luminosities to derive SFRs of the HAEs instead of \ha\ luminosities, we also find the existence of HAEs above the MS at the faint end. Our results suggest that while the majority of massive galaxies are already settled in a secular evolution phase and are thus found on the MS, some less massive galaxies are in a starburst phase and they are significantly up-scattered from the MS. This may be consistent with the down-sizing scenario of mass-dependent galaxy evolution \citep[e.g.,][]{Cowie1996,Bundy2006,Muzzin2013}, or since they are located in a dense proto-cluster, they may be experiencing some influences from the surrounding environment such as galaxy-galaxy interactions. We do not know, however, if this trend is seen only in high density regions or it is a common feature of less massive SFGs (traced by \ha) irrespective of environment. The SED fitting described in \S \ref{sec:MS} indicates that the youngest age of $<10^8$ yr is preferred for galaxies with stellar mass less than $10^{9}$ \Msun, which is again consistent with the less massive HAEs having sSFR of $>10^{-8}$ yr$^{-1}$. This also supports our interpretation that they are young, star-bursting galaxies during the vigorous formation/assembly epoch of a rich galaxy cluster. \citet{Cooke2014} show the lack of galaxies with stellar mass less than $\sim10^{10}$ \Msun\ in a proto-cluster at $z=2.49$, and argue that it is possibly due to either a large dust extinction of less massive galaxies or the earlier formation of massive galaxies. However, our results show that there are SFGs on the MS down to stellar mass of $10^{9.3}$ \Msun\ and that even at lower mass bin there are SFGs with SFRs comparable to those of more massive galaxies with $10^{10}$ \Msun, which are not in agreement with the results by \citet{Cooke2014}. Compared to the USS1558, the proto-cluster discussed in \citet{Cooke2014} is not very rich, although it shows some overdensity in contrast to the general fields. Therefore, the discrepant result between this {\it Letter} and \citet{Cooke2014} could be due to the intrinsic diversity of the properties of proto-clusters at $z\sim2.5$. However, to address this issue, it is essential for us to investigate a much larger sample of proto-clusters. The existence of the less massive HAEs with $<10^{9.3}$ \Msun\ up-scattered above the MS may imply that a scatter around the MS increases at lower stellar masses. Diversity of star-formation history in early phase of galaxy evolution and/or sensitivity to the fluctuation of starburst activity at short time scales in individual \HII\ regions could cause the increased scatter. Another possible remaining issue is a metallicity dependence of the \ha\ luminosity \citep[e.g.,][]{Bicker2005,Dopita2006}. A lower stellar metallicity would result in a higher stellar temperature, and thus the larger number of ionizing photons. Therefore, the SFRs for less massive galaxies can be overestimated due to the metallicity effect, if they follow the mass-metallicity relation \citep[e.g.,][]{Shimakawa2015a}. These are areas of research for future papers.

1607.04040

1607

1607.06874_arXiv.txt

Polarized scattering in planetary atmospheres is computed in the context of exoplanets. The problem of polarized radiative transfer is solved for a general case of absorption and scattering, while Rayleigh and Mie polarized scattering are considered as most relevant examples. We show that (1) relative contributions of single and multiple scattering depend on the stellar irradiation and opacities in the planetary atmosphere; (2) cloud (particle) physical parameters can be deduced from the wavelength-dependent measurements of the continuum polarization and from a differential analysis of molecular band absorption; (3) polarized scattering in molecular bands increases the reliability of their detections in exoplanets; (4) photosynthetic life can be detected on other planets in visible polarized spectra with high sensitivity. These examples demonstrate the power of spectropolarimetry for exoplanetary research and for searching for life in the universe.

\label{sec:prt} Radiative processes in planetary atmospheres is a classical subject, simply for the reason that we live in one. Extensive theoretical studies were carried out during the second half of the twentieth century by such giants as Sobolev \cite{Sobolev1956} and Chandrasekhar \cite{Chandra1960} as well as the renown radiative transfer school at the Saint Petersburg (Leningrad) University \cite{Nagirner2016}. Most recently, physics of planetary atmospheres has become one of the most acclaimed subjects because of applications for Earth climate studies and the detection of a large variety of extrasolar planets. This paper provides the theoretical basis for studying atmospheres of exoplanets using techniques of spectropolarimetry available to us. In particular, using molecular band and continuum spectropolarimetry one can reveal the composition of the gaseous atmosphere, particle layers (clouds, hazes, etc.) and the planetary surface, including the land, water, and life. Modeling these cases is described in this paper. We start from solving a self-consistent radiative transfer problem for polarized scattering in a planetary atmosphere illuminated by a host star. We solve this problem under the following assumptions: 1) the atmosphere is plane-parallel and static; 2) the planet is spherically symmetric; 3) stellar radiation can enter the planetary atmosphere from different angles and can be polarized; 4) an incoming photon is either absorbed or scattered according to opacities in the atmosphere; 5) an absorbed photon does not alter the atmosphere (model atmosphere includes thermodynamics effects of irradiation); 6) photons can be scattered multiple times until they escape the atmosphere. These assumptions expand those in \cite{FluriBerd2010}, namely that multiple scattering is allowed, stellar irradiation can be polarized and vary with an incident angle, and the planetary atmosphere can be inhomogeneous in both longitude and latitude. Then, the radiative transfer equation for the Stokes vector ${\bf I}=(I,Q,U,V)^\mathrm{T}$ of scattered polarized radiation of a given frequency (omitted for clarity) towards ($\mu=\cos\theta$, $\varphi$) is \begin{equation}\label{eq:rt} \mu\frac{d\bf{I}(\tau,\mu,\varphi)}{d\tau} = \bf{I}(\tau,\mu,\varphi) - \bf{S}(\tau,\mu,\varphi) \end{equation} with the total source function \begin{equation}\label{eq:sourcef} \bf{S}(\tau,\mu,\varphi) = \frac{\kappa(\tau)\bf{B}(\tau)+\sigma(\tau)\bf{S}_\mathrm{sc}(\tau,\mu,\varphi)}{\kappa(\tau)+\sigma(\tau)} \rm \ , \end{equation} where $\kappa$ and $\sigma$ are absorption and scattering opacities, $\bf{S}_\mathrm{sc}$ and $\bf{B}$ are the scattering source function and the unpolarized thermal emission, respectively, and $\tau$ is the optical depth in the atmosphere with $\tau=0$ at the top. The formal solution of Eq.~(\ref{eq:rt}) is (e.g., \cite{Sobolev1956}) \begin{equation}\label{eq:rt_formal} \bf{I}(\tau,\mu,\varphi) = \bf{I}(\tau_*,\mu,\varphi)e^{-(\tau_*-\tau)/\mu} + \int_{\tau}^{\tau_*} \bf{S}(\tau',\mu,\varphi)e^{-(\tau'-\tau)/\mu}\frac{\mathrm{d}\tau'}{\mu} \rm \ , \end{equation} where $\tau_*$ is either the optical depth at the bottom of the atmosphere for the Stokes vector $\bf{I}^+(\tau,\mu,\varphi)$ coming from the bottom to the top ($\theta < \pi/2$) or the optical depth at the top of the atmosphere ($\tau_*=0$) for the Stokes vector $\bf{I}^-(\tau,\mu,\varphi)$ coming from the top to the bottom ($\theta > \pi/2$). The scattering source function $\bf{S}_\mathrm{sc}$ is expressed via the scattering phase matrix $\bf{\hat{P}}(\mu,\mu';\varphi,\varphi')$, depending on the directions of the incident ($\mu'$, $\varphi'$) and scattered ($\mu$, $\varphi$) light: \begin{equation}\label{eq:sourcef_sc} \bf{S}_\mathrm{sc}(\tau,\mu,\varphi) = \int \bf{\hat{P}}(\mu,\mu';\varphi,\varphi')\bf{I}(\tau,\mu',\varphi')\frac{d\Omega'}{4\pi} \rm \ . \end{equation} It has contributions from scattering both incident stellar light and intrinsic thermal emission. Their relative contributions depend on the frequency. For instance, for Rayleigh scattering the intensity of the thermal emission of a relatively cold planet in the blue part of the spectrum may become negligible compared to that of the scattered stellar light. The phase matrix $\bf{\hat{P}}(\mu,\mu';\varphi,\varphi')$ is a $4\times4$ matrix with six independent parameters for scattering cases on particles with a symmetry \cite{HansenTravis1974}. In this paper we employ the Rayleigh and Mie scattering phase matrices but our formalism is valid for other phase functions too. The Stokes vector of the light emerging from the planetary atmosphere ${\bf I}(0,\mu,\varphi)$ is obtained by integrating iteratively Equations (\ref{eq:sourcef}) and (\ref{eq:rt_formal}) for a given vertical distribution of the temperature and opacity in a planetary atmosphere. Boundary conditions are defined by stellar irradiation at the top, planetary thermal radiation at the bottom, and (if present) reflection from the planetary surface. Stellar irradiation can be polarized, but the planetary thermal radiation is unpolarized. In particular, stellar limb darkening and linear polarization due to scattering in the stellar atmosphere \cite{KostoBerd2015,Kosto2016} can be taken into account, including the influence of dark spots on the stellar surface \cite{Kosto2015}. This effect is not very large but may be important for cooler stars with large spots and planets on very short-period orbits (when the stellar radiation incident angle noticeably vary depending on the stellar limb angle). Also, stellar magnetic fields causing polarization in stellar line profiles due to the Zeeman effect can be included for given atomic and molecular lines \cite{Berd2003}. This effect is only important for high-resolution spectropolarimetry which is not yet possible for exoplanets. Depending on the structure of the phase matrix and the boundary conditions, the equations are solved for all or a fewer Stokes vector components. Normally it takes 3--7 iterations to achieve a required accuracy. The radiation flux is then obtained by integrating the Stokes vector over the illuminated planetary surface with a coordinate grid ($6^\circ\times6^\circ$) on the planetary surface for a given orbital phase angle as described in \cite{FluriBerd2010}. Our model includes the following opacity sources: (1) Rayleigh scattering on H I, H$_2$, He I, H$_2$O, CO, CH$_4$ and other relevant molecules, Thomson scattering on electrons, and Mie scattering on spherical particles with a given size distribution, with all scattering species contributing to the continuum polarization, (2) absorption in the continuum due to free-free and bound-free transitions of H I, He I, H$^-$, H$_2$$^+$, H$_2$$^-$, He$^-$, metal ionization, and collision-induced absorption (CIA) by H$_2$--H$_2$; (3) absorption and scattering in atomic and molecular lines for particular frequencies where they contribute. Number densities of the relevant species are calculated with a chemical equilibrium code described in \cite{Berd2003}. Here we employ model atmospheres from \cite{Allard2001} and \cite{Witte2009} for stellar and planetary atmospheres, respectively, according to their effective temperatures (T$_{\rm eff}$). This is appropriate for illustrating radiative transfer effects discussed in Section~\ref{sec:res} and applicable for the case of highly irradiated hot Jupiters and substellar components. In particular, a model atmosphere of a hot Jupiter has to match the infrared thermal radiation of the planet originating in deeper layers, while upper layers contributing to the optical radiation are completely dominated by the incident stellar radiation. Planetary atmosphere models with specific chemical compositions and temperature-pressure (TP) structures can be also employed. For instance, the planetary atmosphere can be inhomogeneous with the vertical composition and TP-structure varying with latitude and longitude.

1607.06874

1607

1607.01386_arXiv.txt

At translinear scales, the log power spectrum captures significantly more cosmological information than the standard power spectrum. At high wavenumbers $k$, the Fisher information in the standard power spectrum $P(k)$ fails to increase in proportion to $k$ in part due to correlations between large- and small-scale modes. As a result, $P(k)$ suffers from an information plateau on these translinear scales, so that analysis with the standard power spectrum cannot access the information contained in these small-scale modes. The log power spectrum $P_A(k)$, on the other hand, captures the majority of this otherwise lost information. Until now there has been no means of predicting the amplitude of the log power spectrum apart from cataloging the results of simulations. We here present a cosmology-independent prescription for the log power spectrum; this prescription displays accuracy comparable to that of \citet{Smith_et_al}, over a range of redshifts and smoothing scales, and for wavenumbers up to $1.5h$ Mpc$^{-1}$.

\label{sec-intro} Three-dimensional galaxy surveys can potentially yield significant gains in our understanding of cosmological parameters. Realizing this potential, however, depends on our ability to extract the cosmological information inherent in translinear modes. The power spectrum $P(k)$ is the standard means of elucidating such information (e.g., \citealp{Peebles1980,BaumgartFry1991}); and if one is analyzing a Gaussian field, this statistic exhausts the field's information\footnote{We use the term ``information'' as a shorthand for the Fisher information content \citep{Fisher1925} of the probability density function of the matter fluctuation field (e.g., \citealp{Tegmark1997}).}. Given that inflation models typically predict a high degree of Gaussianity for primordial fluctuations (e.g., \citealp{Bardeen1986, BondEfstathiou1987}); given that the Cosmic Microwave Background (CMB) indeed displays a high degree of Gaussianity; and given that the evolution of the power spectrum at small wavenumbers is essentially linear, it follows that the matter power spectrum is an effective summary statistic for capturing the large-scale information in three-dimensional surveys. On smaller (translinear) scales, however, gravitational amplification of the original fluctuations alters the dark matter field to a distinctly non-Gaussian distribution \citep{FryPeebles1978, Sharp1984, Szapudi1992, Bouchet1993, Gaztanaga1994}. As a result of this nonlinear evolution, the distribution develops a long non-Gaussian tail; this tail produces large cosmic variance, since stochastic occurrence of massive clusters disproportionately affects the power spectrum on such scales \citep{Neyrinck2006}. This increase in cosmic variance involves a corresponding decrease in the information content of the power spectrum. One aspect of this information loss is the rise of correlations between the values of $P(k)$, so that extending the survey to smaller scales only modestly increases the Fisher information in $P(k)$. Through this and other mechanisms, a significant amount of information escapes from the power spectrum, producing a marked information plateau at such wavenumbers \citep{RimesHamilton2005, NeyrinckSzapudi2007, LeePen2008, Carron2011, CarronNeyrinck2012, Wolk2013}. \citet{Repp2015} have shown that for amplitude-like parameters, the power spectrum on translinear scales can contain an order of magnitude less information than it would for a Gaussian field. Since survey forecasts typically assume Gaussianity, they can thus overestimate a survey's effectiveness (and hence its effective volume) by a factor of two or more. Higher order statistics (e.g., $N$-point correlation functions) can access some \citep{Szapudi2009} but not all of \citep{CarronNeyrinck2012, CarronSzapudi2013} this information; these statistics also suffer from difficulties in calculation and interpretation. The log transformation, on the other hand, is particularly attractive as a means of accessing this information \citep{Neyrinck2006, NSS09} in that it emerges naturally under the assumption of linear growth of peculiar velocities \citep{ColesJones1991}. Despite its simple analytic form, \citet{SzapudiKaiser2003} have shown that this transformation is equivalent to an infinite-order loop perturbation theory. In addition, simulations have indicated that the shape of the log power spectrum tracks that of the linear power spectrum up to high wavenumbers \citep{NSS09}, thus reversing the effects of nonlinear evolution. \citet{CarronSzapudi2013} investigate observables which can extract all of the cosmological information inherent in a field. Statistics constructed from such observables (``sufficient statistics'') would, if tractable, be the optimal statistics to use in analyzing matter distributions. In particular, they consider the observable $A$ produced by applying the log transformation to the overdensity field $\delta = \rho/\overline{\rho} - 1$: \begin{equation} A = \ln (1 + \delta). \end{equation} \citeauthor{CarronSzapudi2013} show that the observable $A$ differs from the true optimal observable by a negligible amount, as long as the power spectrum slope is reasonably close to $-1$. Thus, despite the fact that the Universe's matter distribution is only approximately lognormal, the log overdensity $A$ is essentially a sufficient statistic for these fields. One could argue, of course, that in some sense the ``information'' in the fields $\delta$ and $A$ is identical, given the existence of a known invertible mapping between them. However, what \citeauthor{CarronSzapudi2013} (among others) show is that at high wavenumbers, the Fisher information in the log power spectrum $P_A(k)$ is comparable to the information in the field---whereas the power spectrum $P(k)$ contains significantly less information, as would indeed any combination of higher-order statistics of the $\delta$ field. The log density fluctuation $A$, together with its power spectrum $P_A(k)$, is thus an ideal means of completely extracting cosmological information on translinear scales. However, the primary barrier to the use of $P_A(k)$ has been the lack of a theory predicting its value for various cosmologies. This lack leaves simulations as the primary---and computationally expensive---means of exploring $P_A(k)$. In this letter we present a simple prescription for predicting the log power spectrum $P_A(k)$. This prescription contains only one phenomenological parameter, and it provides accuracy comparable to that of \citet{Smith_et_al} as refined by \citet{Takahashi2012}. For the remainder of this letter we use ST to denote this standard prescription of \citeauthor{Smith_et_al} and \citeauthor{Takahashi2012} We organize this letter as follows: Section~\ref{sec:method} outlines the process by which we derive our basic prescription. Section~\ref{sec:accuracy} quantifies the accuracy of our prescription; it also demonstrates that including a slope modulation parameter substantially increases the accuracy. Section~\ref{sec:discussion} discusses potential future refinement to and application of the prescription; and Section~\ref{sec:concl} summarizes this work.

\label{sec:concl} Based on the Millennium Simulation, we have provided a prescription (Equations~\ref{eq:sigA} and \ref{eq:PA}) for calculating the log power spectrum $P_A(k)$. This prescription is accurate to a few percent when one includes a slope modulation parameter $\alpha$ (Equations~\ref{eq:slopemod}--\ref{eq:N}), and this accuracy is comparable to (indeed, better than) that of the \citet{Smith_et_al} prescription for $P(k)$. Our prescription is accurate for redshifts from $z=0$ to at least 2, on smoothing scales from 2 to 16 Mpc$/h$, and to wavenumbers past $1.5h$ Mpc$^{-1}$. It is independent of cosmology, as long as one stays reasonably close to the concordance cosmology. Previous work has conclusively demonstrated the utility of the log power spectrum in constraining cosmological parameters. \citet{WCS2015} show that it contains twice as much cosmological information on parameters such as $\sigma_8$ and $w_0$ (see also \citealp{WCS2015Forecast}). \citet{Repp2015} show that proper accounting for non-Gaussianity (which use of the log power spectrum accomplishes) can increase the effective dark energy figure of merit by a factor of three. Similarly, \citet{Wolk2015} show that this technique can tighten the constraint on neutrino mass by a factor of three. However, to access this information one must precisely predict the log power spectrum $P_A(k)$. In this work we have prescribed a simple means of doing so, and this prescription thus paves the way for a significant increase in the precision of our cosmological knowledge.

1607.01386

1607

1607.00043_arXiv.txt

We evaluate the covariance matrix of the matter power spectrum using perturbation theory up to dominant terms at 1-loop order and compare it to numerical simulations. We decompose the covariance matrix into the disconnected (Gaussian) part, trispectrum from the modes outside the survey (beat coupling or super-sample variance), and trispectrum from the modes inside the survey, and show how the different components contribute to the overall covariance matrix. We find the agreement with the simulations is at a 10\% level up to $k \sim 1 h {\rm Mpc^{-1}}$. We show that all the connected components are dominated by the large-scale modes ($k<0.1 h {\rm Mpc^{-1}}$), regardless of the value of the wavevectors $k,\, k'$ of the covariance matrix, suggesting that one must be careful in applying the jackknife or bootstrap methods to the covariance matrix. We perform an eigenmode decomposition of the connected part of the covariance matrix, showing that at higher $k$ it is dominated by a single eigenmode. The full covariance matrix can be approximated as the disconnected part only, with the connected part being treated as an external nuisance parameter with a known scale dependence, and a known prior on its variance for a given survey volume. Finally, we provide a prescription for how to evaluate the covariance matrix from small box simulations without the need to simulate large volumes.

\label{sec:intro} The distribution of matter in the Universe contains a wealth of information about the energy content of the Universe, its properties, and evolution. Initial distribution is thought to be a Gaussian random field, but as a result of the gravitational instability the tiny fluctuations in the initial matter distribution evolves nonlinearly and produces non-Gaussian correlations. The simplest statistic for data analysis is the two point correlation function, or its Fourier transform the power spectrum. This contains significant cosmological information since it is sensitive to many parameters, and much of the information comes from deeply in the nonlinear regime. The two-point analysis is the prime focus of many cosmological observables like weak lensing (WL) or galaxy clustering, which are the key observable probes in the current and future generation surveys like Dark Energy Survey (DES\footnote{http://www.darkenergysurvey.org}), Dark Energy Spectroscopic Instrument (DESI\footnote{http://desi.lbl.gov}), Large Synoptic Survey Telescope (LSST\footnote{https://www.lsst.org}), Euclid\footnote{http://www.euclid-ec.org} etc. Our ability to extract useful constraints on the cosmological parameters from these surveys depend on our ability to model the statistical properties of the distribution of matter in the Universe, particularly the matter power spectrum, in the non-linear regime ($k > 0.2 h^{-1}{\rm Mpc}$). In recent years, there have been many efforts in providing accurate estimates of the matter power spectrum for various cosmological models and redshifts using perturbation theory \citep{2002PhR...367....1B}, the halo model \citep{2002PhR...372....1C}, simulations \citep{2010ApJ...715..104H,2009ApJ...705..156H,2010ApJ...713.1322L}, and semi-analytic models \citep{2014MNRAS.445.3382M,2015PhRvD..91l3516S}. The current precision from simulations is at 1\%, and can easily be improved further with simulations if necessary. There are additional complications such as the baryonic effects \citep{2011MNRAS.415.3649V, 2011MNRAS.417.2020S, 2013MNRAS.434..148S, 2014arXiv1410.6826M}, which redistribute the matter within the halo centers and change the power spectrum at high $k$, and it is likely that WL will not be able to extract much reliable information at high $k$ (or $l$). We will not model these processes here and focus on the dark matter part only. While the matter power spectrum predictions, in the absence of baryonic effects, are under control, for a complete analysis one also needs its covariance matrix. The covariance matrix of the matter power spectrum is important in order to perform any statistical inference analysis on the cosmological data. Therefore, future surveys would require an accurate estimation of the covariance matrix of the matter power spectrum in order to perform the cosmological parameters estimation. An accurate quantification of the covariance matrix is crucial in order to derive constraints on the cosmological parameters using observables modeled on the matter power spectrum. Failing to do so will mislead the interpretation of the data. In the linear regime, computing the covariance matrix is reasonably straight-forward, even if it can be computationally expensive for surveys with complicated masks. Due to the non-linear evolution of the matter density field, the Fourier modes become correlated, and this generates a non-Gaussian covariance matrix \citep{1999ApJ...527....1S, 2001ApJ...554...67H,2001ApJ...554...56C}, which we also call the connected part. It affects both diagonal and non-diagonal elements of the covariance matrix. Its characterization has been developed in terms of the physically motivated halo model \citep{2013PhRvD..87l3504T,2014MNRAS.445.3382M}, or numerical simulations \citep{2009ApJ...700..479T,2011ApJ...734...76S,2012MNRAS.423.2288H,2013PhRvD..88f3537D,2014PhRvD..89h3519L, 2015MNRAS.446.1756B,2015arXiv151205383B}. Another approach is using perturbation theory (PT), which is the focus of the work here. Our approach is to identify the perturbative terms that dominate the covariance matrix, and PT can be used to understand better the structure of the covariance matrix. This paper is organized as follows: In section \ref{sec:newmodel}, we review the approach and describe the perturbation theory up to partial 1-loop calculations for the full covariance matrix of the matter power spectrum. In section \ref{sec:sims}, we outline a comparison of our model with two sets of cosmological simulations, with and without SSC term. In section \ref{sec:decomposition}, we decompose the covariance matrix based on simulations and the analytic model into a vector that fully describes the full covariance matrix, with and without SSC term, using a principle component analysis. In section \ref{sec:fisher}, we study the degeneracies between covariance matrix model and other cosmological parameters and perform a Fisher information matrix analysis. We also address the question of how to evaluate covariance matrix from small box simulations. Finally we conclude with a discussion in section \ref{sec:discussion}. During the completion of our work an independent analysis including all PT terms up to 1-loop has been presented here \cite{2015arXiv151207630B}, with results comparable to ours.

\label{sec:discussion} In this paper, we developed a perturbative model for the covariance matrix of the power spectrum, going up to 1-loop in PT, but without including all the 1-loop terms (see \cite{2015arXiv151207630B} for the full 1-loop calculation). In addition, we go beyond 1-loop by including the nonlinear damping of response functions \citep{2014arXiv1411.2970N}, which improves our results at higher $k$. Overall, our approach predicts the results from simulations to about ten percent accuracy in the quasi-linear regime. The largest contribution to the covariance is beat coupling or super-sample covariance (SSC), which arises due to the coupling of the modes to the ones which are larger than the survey scale and which has been extensively studied previously \citep{2014PhRvD..89h3519L,2014PhRvD..90j3530L}. This effect comes from the modes outside the survey and cannot be captured by the jackknife or bootstrap methods, which subdivide the full volume into many smaller volumes. In fact, these methods overestimate the covariance terms because they generate SSC from the smaller regions. We propose an alternative approach which removes this problem. On large scales, the second largest contribution comes from tree-level terms from modes inside the survey, while on smaller scales 1-loop terms dominate. The dominant 1-loop term, and the only one we include, comes from the sampling variance fluctuations of large-scale modes within the survey volume, which induce a coherent response by small-scale modes which we model using 1-loop power spectrum. When we allow for a damped nonlinear response to modes that are highly nonlinear \citep{2014arXiv1411.2970N}, we obtain some additional improvements, especially at lower redshifts. Damping does not affect results at higher redshifts because all of the modes that have significant sampling fluctuations are linear. We explore the structure of the covariance matrix and find that its non-Gaussian part is strongly dominated by a single eigenmode. This suggests that the non-Gaussian response has always the same shape, only its amplitude varies. We analyze the probability distribution of this amplitude and find that it is close to a Gaussian. Thus, the convergence rate of covariance matrix simulations scales as a Gaussian, with $(2/N)^{1/2}$ giving the relative error after $N$ simulations. One possible alternative approach is to ignore the non-Gaussian covariance in the analysis and include instead the eigenvector response as a fictitious external parameter in the analysis. This parameter can in principle be determined from the data itself, but it is quite degenerate with other cosmological parameters. While the analysis given in this paper provides several important insights into the nature of the covariance matrix of the two-point correlators, our results cannot be directly applied to the data. For weak lensing observations, which are measuring the dark matter correlations, one needs to perform the projection along the line of sight. This leads to a decorrelation of the non-Gaussian covariance because different $l$ correspond to different effective redshift, and hence different effective volume. This differs from our analysis where all modes are maximally correlated because they are coupled to the same long wavelength modes within the given volume. So far this has only been addressed in the context of the halo or SSC model \citep{2014PhRvD..90l3523S,2016arXiv160105779K}. The 3-d analysis performed here is more likely to be of immediate application to the galaxy clustering, but this would require adding biasing and redshift space distortions to the model, and it is unclear whether successful analytic models can be built. Nevertheless, the decomposition of the covariance matrix into the three components, disconnected, connected but generated from modes outside the survey, and connected and generated from modes inside the survey, allows one in principle to build the full covariance matrix from relatively small simulation volumes, and possibly even from the data itself, by subdividing into sub-volumes and determining each of the three components separately. It is worth pursuing this further to determine the optimal approach that delivers the highest accuracy with this technique.

1607.00043

1607

1607.07725_arXiv.txt

We propose that $\gamma$-rays in blazars can be produced during encounters of relativistic blobs of plasma with radiation field produced by luminous stars within (or close to) the jet. The blob is expected to contain relativistic electrons which comptonize stellar radiation to the GeV-TeV energies. Produced $\gamma$-rays can initiate the Inverse Compton $e^\pm$ pair cascade in the stellar radiation. We propose that such a scenario can be responsible for the appearance of the so-called orphan $\gamma$-ray flares. We show that the relativistic blob/luminous star collision model can explain the appearance of the extreme orphan $\gamma$-ray flare observed in the GeV and sub-TeV energy range from the flat spectrum radio quasar PKS\,1222+21.

It is widely accepted that the non-thermal radiation in blazars is produced within a relativistic jet directed towards the observer. However, the mechanism of energization of particles and their radiation is not fully understood. Recently, simple scenarios exploiting collisions of compact objects with the jet plasma reached some attention. In fact, many different types of compact objects (e.g. stars, clouds or even globular clusters, fragments of supernova remnants or pulsar wind nebulae) can be immersed within the jet, forming an obstacle for the jet plasma. It has been proposed that particles, accelerated on the shocks appearing as a result of collisions of the jet plasma with stellar winds, can produce high energy $\gamma$-rays (see e.g. \citealp{bp97,ba10, br12, ar13, wy14, bb15, br15, de16}). In this work, we consider another encounter scenario in which relativistic particles, already present in the fast moving blob, interact with the dense radiation of a star within the jet. Relativistic leptons, which are isotropic in the blob, comptonize stellar radiation to GeV-TeV energies. The radiation field of the star is strong enough that $\gamma$-rays produced in the Inverse Compton process can initiate the $e^\pm$ pair cascade in the stellar radiation. To study such a scenario we developed a dedicated Monte Carlo code tracking an electromagnetic cascade in the anisotropic radiation field of a star. We calculate the $\gamma$-ray spectra emerging from such process and discuss their features. As an example, we show that the $\gamma$-ray spectrum and light curve observed from a distant Flat Spectrum Radio Quasar (FSRQ) PKS\,1222+21 during its 2010 orphan $\gamma$-ray flare might be explained in terms of such a scenario. We propose that blobs of relativistic plasma, moving in jets of active galactic nuclei (AGNs), can encounter from time to time the luminous stars which form a quasi-spherical halo around the central super-massive black hole (see Fig.~\ref{fig1}). For example, in the nearby active radio galaxy Cen~A, the current star formation rate ($\sim$2 M$_\odot$ yr$^{-1}$) should result in the production of $(6 -12)\times 10^7$ M$_\odot$ of young stars during the duty cycle of the Cen A active nucleus. We can thus expect $\sim$3$\times 10^5$ stars with masses above 20 M$_\odot$ \citep{wy14}. A large number of these stars should be immersed in the jet pointing towards the observer. In the case of a simple conical jet, the number of stars is estimated to be of the order of $\sim$83$\gamma_{30}^{-2}$, where we assumed that the jet opening angle is determined by the Lorentz factor of the flow $\theta \sim 1/\gamma_{\rm b}$ and $\gamma_{\rm b}=30\times \gamma_{30}$. In the case of a parabolic jet, the number of stars depends in a more complicated way on the parameters of the jet but is expected to be also significant since the perpendicular extend of the jet on a parsec distance scale from its base are of similar order (see estimates in Sect.~\ref{sec3}). The Cen~A nucleus is also immersed in a quasi-spherical bulge of late type stars with the number estimated as $\sim 8\times 10^8$ \citep{wy14}. The total number of luminous red giants can be as large as $\sim$10$^6$, if only $\sim$10$^{-3}$ of these bulge stars are in the red giant phase. Therefore, the interaction of a relativistic blob with the radiation field of such luminous stars within a jet of an active galaxy seems to be quite likely. In Section~\ref{sec3} we explain the basic properties of the blob-star interaction model. In Section~\ref{sec4} we apply such scenario to the extreme flare observed from PKS\,1222+21. \kom{In Section~\ref{sec5}, we summarize our conclusions.}

\label{sec5} We propose that blobs of relativistic plasma, moving in jets of AGNs, can encounter from time to time luminous stars which form a quasi-spherical halo around the central super-massive black hole. The transit time of a star through a conical jet can be estimated as $T_{\rm transit} = R_{\bot}/v_\star\approx 0.45\gamma_{30}^{-1}l_{-1}^{3/2}/M_9^{1/2}$~yr, where the velocity of the star is $v_\star = \sqrt{GM_{\rm BH}/l}\approx 7\times 10^8 (M_9/l_{-1})^{1/2}$ cm s$^{-1}$, $M_{\rm BH} = 10^9M_9$ M$_\odot$ is the black hole mass in units of the Solar mass, and $G$ is the gravitational constant. \kom{At parsec scale distance from the jet base the jet transit time is similar for a conical and a parabolic, M87-like, jet.} Therefore, we do not expect significant effects of the jet shape on our final results. The star passing the inner jet at parsec distance from the SMBH can be responsible for a sequence of outbursts as different blobs reach its position. Such high activity period can last from months to years. In fact, a single blob may be also responsible for the observed sequence of flares (e.g. as observed in Mrk 421, \citealp{bu96} or PKS\,2155-304, \citealp{ah07}), if it meets on its path a few stars within a jet. Luminous stars are also characterised by stellar winds which pressure can balance the pressure of the blob plasma. As a result a shock structure appears around the star. The shock radius around the star in a conical jet can be determined from (see e.g. \citealp{bp97}), $R_{\rm sh}^\star\approx 1.9\times 10^{12}(M_{-5}v_3)^{1/2}\gamma_{30}^{-1}l_{-1}/L_{46}^{1/2}$ cm, where $\dot{M} = 10^{-5}M_{-5}M_\odot$ and $v = 10^3v_3$ km s$^{-1}$ are the mass loss rate and the velocity of the stellar wind, and $L_{\rm b} = 10^{46}L_{46}$ erg s$^{-1}$ is the power of the blob in the observer's frame. Since the surface area of this shock, $\sim \pi R_{\rm sh}^2$, is typically much smaller than the surface area of the considered blob, the presence of the shock should not essentially influence presented above calculations of the $\gamma$-ray emission from the blob. Our model predicts the appearance of flares which emission is limited mostly to the $\gamma$-ray energy range. Therefore, it can provide the mechanism for the appearance of the so-called orphan $\gamma$-ray flares observed occasionally from relatively nearby AGNs of the BL Lac type (e.g.~ 1ES 1959+650, \citealp{kr04} or Mrk 421, \citealp{fr15}) but also from the FSRQs (e.g. PKS\,1222+21, \citealp{al11}). In order to show applicability of this scenario we model the case of the emission from the very strong $\gamma$-ray flare observed from the distant FSRQ PKS\,1222+21 in 2010. Our model is able to reproduce the observed light curve and spectral energy distribution measured by the {\it Fermi}-LAT and MAGIC, although the energy requirements are quite demanding. Note that in the case of blazars with orphan flares which are about an order of magnitude closer to the observer (i.e. 1ES 1959+650 and Mrk 421), this energy requirements should be about two orders of magnitude lower or the Lorentz factors of the blob might be closer to values typically reported from radio observations.

1607.07725

1607

1607.03937_arXiv.txt

Many exoplanetary systems containing hot Jupiters (HJs) exhibit significant misalignment between the spin axes of the host stars and the orbital angular momentum axes of the planets (``spin-orbit misalignment''). High-eccentricity migration involving Lidov-Kozai oscillations of the planet's orbit induced by a distant perturber is a possible channel for producing such misaligned HJ systems. Previous works have shown that the dynamical evolution of the stellar spin axis during the high-$e$ migration plays a dominant role in generating the observed spin-orbit misalignment. Numerical studies have also revealed various patterns of the evolution of the stellar spin axis leading to the final misalignment. Here we develop an analytic theory to elucidate the evolution of spin-orbit misalignment during the Lidov-Kozai migration of planets in stellar binaries. Secular spin-orbit resonances play a key role in the misalignment evolution. We include the effects of short-range forces and tidal dissipation, and categorize the different possible paths to spin-orbit misalignment as a function of various physical parameters (e.g. planet mass and stellar rotation period). We identify five distinct spin-orbit evolution paths and outcomes, only two of which are capable of producing retrograde orbits. We show that these paths to misalignment and the outcomes depend only on two dimensionless parameters, which compare the stellar spin precession frequency with the rate of change of the planet's orbital axis, and the Lidov-Kozai oscillation frequency. Our analysis reveals a number of novel phenomena for the stellar spin evolution, ranging from bifurcation, adiabatic advection, to fully chaotic evolution of spin-orbit angles.

The discovery of the misalignment between the orbital axes of hot Jupiters (HJs; giant planets with orbital periods of $\sim 3$~days) and the spin axes of their host stars (e.g. Hebrard et al. 2008, 2010; Narita et al. 2009; Triaud et al. 2010; Winn et al. 2009; Albrecht et al. 2012; Winn \& Fabrycky 2015) continues to pose a significant puzzle. While primordial disk misalignment (with respect to the stellar spin) is a possible explanation (e.g. Bate et al. 2010; Lai et al.~2011; Foucart \& Lai 2011; Batygin 2012; Batygin \& Adams 2013; Lai 2014; Spalding \& Batygin 2014), it is likely that a significant fraction of HJs and the associated spin-orbit misalignments are produced by dynamical means involving multi-planet interactions or planet-binary interactions. Lidov-Kozai (LK) oscillations (Lidov 1962, Kozai 1962) induced by external stellar or planetary companions -- one of the proposed channels of hot Jupiter formation (e.g. Wu \& Murray 2003; Fabrycky \& Tremaine 2007; Correia et al 2011; Beaug{\'e} \& Nesvorn{\'y} 2012; Naoz et al. 2012; Petrovich 2015; Anderson, Storch \& Lai 2016; Mu{\~n}oz, Lai \& Liu 2016; Petrovich \& Tremaine 2016) -- provide a natural means of generating spin-orbit misalignment. Lidov-Kozai oscillations occur when the proto-HJ's host star has a binary (or external planetary) companion. A proto-HJ is assumed to form at several AUs from its host. If its orbital axis is sufficiently misaligned relative to the outer binary axis, its orbit undergoes large correlated variations in eccentricity and inclination. If the misalignment is substantial, very high eccentricities (in excess of $0.95$) can be attained; during these high-eccentricity phases, tidal dissipation at periastron brings the planet close to its host, eventually creating a HJ. Since the host star of a giant planet generally has an appreciable rotation (with rotation period ranging from a few days to 30~days) and is oblate, significant coupling can exist between the orbital dynamics of the proto-HJ and the dynamics of the stellar spin axis. This coupling is vital in determining the final spin-orbit misalignments of these systems. Indeed, in Storch, Anderson \& Lai (2014; hereafter SAL14) we showed that the evolution of the stellar spin axis, driven by the quasi-periodic changes in the planet orbit, can be very complex and even chaotic, and this evolution depends sensitively on the planet mass and stellar rotation period. Subsequently, in Storch \& Lai (2015; hereafter SL15), we studied the origin of the chaotic behavior (in terms of secular spin-orbit resonances and their overlaps) by analysing the non-dissipative (i.e. no tidal dissipation) ``stellar spin + planet + binary'' system and considering the regime in which the stellar spin precession rate was much higher than the LK oscillation frequency (the ``adiabatic'' regime). In Anderson, Storch \& Lai (2016; hereafter ASL16), we conducted a comprehensive population synthesis study of HJ formation via LK migration in stellar binaries, including all relevant physical effects (the octupole potential from the binary companion, various short-range forces, tidal dissipation, and stellar spin-down due to magnetic braking). In particular, our extensive Monte-Carlo experiments (see Section 4 of ASL16) revealed various paths of spin-orbit evolution during LK migration. In the present work, we develop an analytic theory to understand the evolution of spin-orbit misalignment during LK migration in stellar binaries. We extend the analysis of SL15 to the non-adiabatic regime (i.e. the regime in which the star precesses slowly). We account for various non-ideal effects (such as periastron advance due to General Relativity and planet oblateness) and include tidal dissipation in our ``stellar spin + planet + binary'' system. Our goal is to provide theoretical explanations for the various paths to spin-orbit misalignment that LK oscillations can induce and to shed light on how the final spin-orbit misalignments of HJs are achieved during LK migration. Although our focus is on stellar binary-induced LK migration, most of our results can be adapted to the planet-induced LK migration scenario. This paper is organized as follows. In Section 2, we briefly review our previous work and introduce the most important concepts and equations in LK-driven spin-orbit dynamics. In Section 3, we discuss the effect that short-range forces have on the spin-orbit dynamics. In Section 4, we examine the different regimes of non-dissipative spin dynamics. In Sections 5 and 6, we include tidal dissipation and study the various paths toward misalignment during LK migration. We identify the key paremeters that determine different behaviors of the misalignment evolution. We discuss the limitations and uncertainties of our work in Section 7 and summarize our key findings in Section 8. Readers who are less interested in the technical details can go to Section 5.1 for a description of the five different spin-orbit evolution paths (with more detailed explanations in Sections 5.2-5.6), and Section 8 for a brief summary.

\subsection{Complications due to Spin Feedback on Orbit} As explained in Section \ref{SRFsec}, a major assumption in the analysis laid out in this paper, as well as in SL15, is the omission of the extra precession the planet's orbit experiences due to perturbation from the stellar quadrupole. This omission enables us to considerably simplify the spin dynamics problem by reducing it to a 1D Hamiltonian system. Stellar feedback on planet's orbit becomes important if the host star has nearly as much, or more, angular momentum as the planet's orbit (see Eq.~\ref{eq:SoverL}). Thus, one may question whether our analysis is truly applicable to Jupiter-mass (as opposed to heavier) planets and rapidly rotating stars. However, in ASL16 we have run a comprehensive suite of numerical simulations including stellar feedback on the orbit and all other important effects. We found that, while under certain conditions the bimodal and stationary adiabatic behaviors expected to dominate for Jupiter-mass planets can be disrupted, on the whole, bimodality remains nearly ubiquitous. We conclude that our classification of the different misalignment evolution modes and outcomes is generally valid. \subsection{Stellar Spindown} In this work, we have assumed that the stellar spin remains constant throughout the proto-HJ's tidal decay and circularization. In reality, solar-type stars experience significant spindown due to magnetic winds. This spindown acts to temporarily reduce the strength of the coupling between the stellar spin and the planet orbit, decreasing both $N_\max$ and $\bar{\mathcal{A}}$. However, it is still the initial values of these parameters ($N_{\rm max,0}$ and $\bar{\mathcal{A}}_0$) that set the qualitative behavior of the system. Some differences in the final value of the spin-orbit misalignment angle $\theta_{\rm sl,f}$ due to stellar spindown could be expected, but the regime classification and predictive power should remain unchanged. \subsection{Primordial Misalignment} In this paper we have focused on stellar spin dynamics in systems that have no initial spin-orbit misalignment. However, since several ways of generating primordial misalignment have been proposed (see references in Section 1), the dynamics of initially misaligned systems are of potential interest. Although a thorough exploration of the dynamics of initially misaligned systems is beyond the scope of this work, we believe the ideas developed in this paper, particularly the importance of the parameters $\abar_0$ and $N_{\rm max,0}$, are still applicable to initially misaligned systems. The fate of such systems should still be determined by the initial phase space of the system and where in that phase space the system is initialized. For example, for trajectories starting anywhere inside the $N=0$ center island (see Fig.~\ref{samplephasespaces}), the final outcome should be bimodal, just as for initially aligned systems; the only difference is that, for an initially misaligned system, the peaks of the bimodal distribution of the final misalignment angle would lie closer to $90^\circ$ due to the smaller initial area of the trajectory. In Fig.~$26$ of ASL16 we have demonstrated that this is the case. Thus, the frameworks developed in this paper for the special case of initially aligned systems can be easily generalized to systems with arbitrary initial misalignments and phases.

1607.03937

1607

1607.01874_arXiv.txt

The Ultraviolet Imaging Telescope (UVIT) was launched as part of the multi-wavelength Indian \astrosat mission on 28 September, 2015 into a low Earth orbit. A 6-month performance verification (PV) phase ended in March 2016, and the instrument is now in the general observing phase. UVIT operates in three channels: visible, near-ultraviolet (NUV) and far-ultraviolet (FUV), each with a choice of broad and narrow band filters, and has NUV and FUV gratings for low-resolution spectroscopy. We have written a software package (\jude) to convert the Level~1 data from UVIT into scientifically useful photon lists and images. The routines are written in the GNU Data Language (GDL) and are compatible with the IDL software package. We use these programs in our own scientific work, and will continue to update the programs as we gain better understanding of the UVIT instrument and its performance. We have released \jude under an Apache License.

The Ultraviolet Imaging Telescope (UVIT) was first proposed as part of the multi-wavelength \astrosat mission in the late $20^{th}$ century \citep{Pati1998,Pati1999} and was launched on 28 September, 2015 into a low Earth orbit (650 km, $6^{\circ}$ inclination) by an ISRO (Indian Space Research Organization) PSLV (Polar Satellite Launch Vehicle) launcher. The UVIT payload doors were opened on 30 November, 2015 and a 6 month performance verification (PV) phase began, ending in March 2016. The ground calibration of the payload has been discussed by \citet{Postma2011} with in-flight tests by \citet{Subramaniam2016}. There are now a complete set of PV phase observations which have been used to characterize the instrument \citep{Tandon2016} with in-flight calibration and verification done by \citet{Tandon2017} and \citet{Rahna2017}, and data are being released to the observers. However, there are still (at the date of writing) issues with the UVIT pipeline and there are data sets which cannot be processed. More importantly from our viewpoint is that the source code is proprietary and difficult to modify. We have chosen to create an alternative set of routines ({\it JUDE}: Jayant's UVIT Data Explorer) released under an Apache License\footnote{\textit{http://www.apache.org/licenses/LICENSE-2.0}} which may be freely used and modified. \jude begins with the Level 1 data provided by the Indian Space Science Data Centre (ISSDC) and produces photon lists and images suitable for scientific analysis. It should be stressed that this package is not intended to be a replacement for the official UVIT pipeline but is rather a tool to examine the data in more detail. However, the authors are using {\em JUDE} for their own scientific purposes, and are actively maintaining and modifying the software as the instrument characterization improves. \jude is archived at the Astrophysics Source Code Library \citep{Murthy_ascl} and the latest version is available on GitHub (https://github.com/jaymurthy/JUDE). We solicit feedback on its operation and further improvements.

We have designed a set of software routines that begin with Level~1 UVIT data from the ISSDC and produces images suitable for scientific purposes. We have tested the software on all our GT observations, and have produced Level~2 event lists and images for all observations through an automated process. A few observations required manual intervention, primarily due to registration problems. We have used {\em JUDE} to characterize the in-flight performance of UVIT \citep{Rahna2017} and now expect to move on to the scientific exploitation of the data. We have released \jude under the Apache License 2.0,\footnote{https://www.apache.org/licenses/LICENSE-2.0.txt} and it is available on the Astrophysics Source Code Library \citep{Murthy_ascl} and on GitHub (\textit{https://github.com/jaymurthy/JUDE}). We are using data from the UVIT for our science and, therefore, will continue to maintain and improve {\em JUDE} as needed. As a final note, the scientific exploitation of the UVIT data has been limited till now owing to the delays in the UVIT pipeline. We hope that with the release of {\em JUDE} to the astronomical community at large, more and better scientific results from the UVIT will be published. The only valid test of scientific software is if it is used widely, and we will continue to update the routines as we find errors or as they are reported to us. We welcome feature requests to improve the scientific utility of the programs.

1607.01874

1607

1607.06453_arXiv.txt

We study time evolution of rotating, axisymmetric, two dimensional inviscid accretion flows around black holes using a grid based finite difference method. We do not use reflection symmetry on the equatorial plane in order to inspect if the disk along with the centrifugal barrier oscillated vertically. In the inviscid limit, we find that the CENtrifugal pressure supported BOundary Layer (CENBOL) is oscillating vertically, more so, when the specific angular momentum is higher. As a result, the rate of outflow produced from the CENBOL, also oscillates. Indeed, the outflow rates in the upper half and the lower half are found to be anti-correlated. We repeat the exercise for a series of specific angular momentum $\lambda$ of the flow in order to demonstrate effects of the centrifugal force on this interesting behaviour. We find that, as predicted in theoretical models of disks in vertical equilibrium, the CENBOL is produced only when the centrifugal force is significant and more specifically, when $\lambda > 1.5$. Outflow rate itself is found to increase with $\lambda$ as well and so is the oscillation amplitude. The cause of oscillation appears to be due to the interaction among the back flow from the centrifugal barrier, the outflowing winds and the inflow. For low angular momentum, the back flow as well as the oscillation are missing. To our knowledge, this is the first time that such an oscillating solution is found with an well-tested grid based finite difference code and such a solution could be yet another reason of why Quasi-Periodic Oscillations should be observed in black hole candidates which are accreting low angular momentum transonic flows.

In astrophysical context, the process in which diffuse gas or matter is accumulated around a compact object under an influence of gravity is called accretion. The importance of accretion as the source of steady or unsteady emission of radiation was first widely recognized in explaining observations of binary systems, especially X-ray binaries. It is very important to understand hydrodynamic properties of matter in the vicinity of a black hole as the emitted radiation mainly depends on the density, velocity and temperature at each flow element at each instant. Unique inner boundary condition which enables all infalling matter to cross the event horizon with the speed of light $c$ (e.g., Chakrabarti, 1996) also makes all the accretion flows into black holes to be transonic in nature forcing them to have the innermost sonic point in regions of strong gravity, i.e., just outside of the event horizon. Due to centrifugal pressure, inflowing matter slows down and piles up closer to the black hole forming a torus like structure at the inner part of the disc. This torus like structure formed between the centrifugally supported shock and the innermost sonic point outside the horizon is widely known as the Centrifugal pressure supported BOundary Layer or simply CENBOL (Chakrabarti, 1989, hereafter C89; 1999; Molteni, Lanzafame \& Chakrabarti, 1994; MLC94). In the limit of no radial velocity, the CENBOL would have been termed as a thick accretion disk (Paczy\'nski \& Wiita, 1980) as shown by MLC94. Earlier, a large number of numerical simulations of inviscid accretion flows around black holes, have been presented by various research groups (e.g., Hawley, Wilson \& Smarr 1984; Eggum, Coroniti, \& Katz 1985; Chakrabarti \& Molteni, 1993; MLC94; Ryu, Brown, Ostriker \& Loeb 1995; Molteni, Sponholz \& Chakrabarti 1996; Molteni, Ryu \& Chakrabarti, 1996, hereafter, MRC96; Ryu, Molteni \& Chakrabarti, 1997, hereafter RMC97; Igumenshchev, Abramowicz \& Narayan 2000; Chakrabarti, Acharya \& Molteni, 2001, hereafter CAM01; Giri et al., 2010, hereafter GC10). However, these simulations were performed assuming that the flow has an equatorial symmetry, and therefore the flow behavior was studied only on one quadrant, i.e., first quadrant using the standard `reflection' boundary condition on the equatorial plane. In MLC94 and GC10, the results of standing and oscillating shock formations in inviscid flows are presented using Smoothed Particle Hydrodynamics (SPH) method and finite difference method respectively. In GC10, several simulations have been carried out choosing two conserved flow parameters (namely, specific energy ${\it E}$ and specific angular momentum $\lambda$) from the parameter space which provides complete set of solutions of a black hole accretion flow (C89). In order to break the reflection symmetry along equatorial plane, several simulations have been carried out by various groups for both black hole accretion (Molteni et al., 2001, hereafter M01; Chakrabarti, Acharya \& Molteni, hereafter CAM01) as well as stellar wind accretion onto stars (Fryxell \& Taam, 1988 ; Taam \& Fryxell, 1989; Matsuda et al., 1991, 1992). In M01 and CAM01, using SPH, it was shown that an instability can occur in the flow. They also demonstrated that although matter is supplied symmetrically, those instabilities may not remain symmetric with respect to the equatorial plane. Furthermore, there is a strong interaction of the outgoing wind with the incoming flow (M01). However, SPH is known to be dissipative in nature and it is not impossible that in energy conserving schemes one might see that such oscillations are actually disrupting the flow altogether. We therefore extend the work of GC10 where energy is accurately preserved by removing the reflection condition along the equatorial plane. By this procedure, we intend to give answers to the following important questions: (a) Will the accretion flow be symmetric with respect to the equatorial plane? And if so, under what conditions? (b) Will this two quadrant flow have any effect on the formation of the so called `CENBOL'? This question is especially relevant as the CENBOL acts as the Compton cloud (Chakrabarti \& Titarchuk, 1995, hereafter CT95) while explaining the spectral and timing properties of black hole candidates. (c) If the flow symmetry is absent then will the accretion flow remain stable at all or the flow would be violent and disrupted? (d) What will be its effects on outflows which are known to be produced on the CENBOL surface? The plan of our paper is the following: in the next Section, we present the model equations governing the flow. In Section 3, we describe the methodology for our simulations using a grid based finite difference technique, called total variation diminishing (TVD) method which was discovered by Harten (1981). In Section 4, we discuss our simulation results and compare those with earlier simulations. Finally, in Section 5, we make concluding remarks.

In this paper, we presented extensive time dependent numerical simulations of two quadrant accretion flow around a black hole. In earlier studies, such as in MLC94 and GC10, similar type of simulations were carried out in one quadrant using SPH and TVD schemes respectively. In these earlier simulations, our goal was to check if the shocks could be produced in the first place, and if yes, how does the puffed up post-shock region behave in reference of thick accretion flow. It was found that indeed, since the post-shock flow is sub-sonic with low radial velocity and sub-Keplerian angular momentum, it does behave as a thick accretion disk, though without a cusp since radial velocity increases close to the inner edge. This post-shock region (CENBOL) was then used to Comptonize the low energy photons from a Keplerian disk. In the present work, we asked ourselves if the CENBOL really remains symmetric with respect to the equatorial plane. For this we removed the reflection symmetry imposed forcefully in earlier simulations. We inject matter only in first and the fourth quadrants. We find that for lower centrifugal force, the CENBOL remains symmetric, though a vertical oscillation sets it which becomes more and more violent as the specific angular momentum increases. This is superimposed with a horizontal oscillation. We also find that the outflow rate from each of the quadrants independently vary: The quadrant in which the CENBOL is tilted, also has the higher rate. Thus, the rates in the two quadrants are anti-correlated. There are two important aspects which need major discussions: First, the oscillations seen here are in radial and vertical directions. In earlier works, especially in Ryu et al. (1997), it was shown that when the Rankine-Hugoniot relations are not satisfied, the shocks undergo radial oscillations. This is because a transonic accretion flow with a significant amount of angular momentum, has two physical sonic points and the flow is required to have higher entropy in order to pass through the innermost one (Chakrabarti, 1989). This means that even when the conditions of steady shock formation is not satisfied, the flow will pass through the inner sonic point and the shock will be unsteady, moving radially, and searching for an acceptable solution. Another reason of radial oscillation was found by Molteni et al. (1996a) where it was seen that shocks start to oscillate only when the cooling time scale in the post-shock region roughly matches with the infall time scale in that region. Since cooling process is absent in the present simulation, this latter explanation is not relevant in the present context. In presence of viscosity, oscillation due to viscous overstability (Kato, 1978) is well known. But ours is an inviscid flow. Hence this is also not the reason of the oscillation seen in our simulations. In case of slender tori, inertial modes are rapidly excited faster than the dynamical time scale (Blaes, 2006) and causes significant instability (Horak, 2012). However, our post-shock region is advecting and not slender. Perturbations due to the inertial modes are likely to be advected out of the disk when the flow passes through the innermost sonic point. In the present context, we note that the disk instability is high only when the angular momentum is very large. Two important physical processes are triggered by angular momentum: (i) infalling matter hits the centrifugal barrier (defined by the location where the centrifugal forces is similar to the gravitational force) and bounces back near the equatorial plane. This flow confronts the incoming matter and two turbulence cells of opposite vorticity are generated, one above and the other below the equatorial plane. (ii) Centrifugal pressure driven winds are formed which also flow outwards (in between the centrifugal barrier and the so-called funnel wall, see, Molteni et al. 1996b), confronting the incoming flow away from the equatorial plane. This interface is therefore susceptible to Kelvin-Helmholtz instability. In all the cases we ran, we found that for very low angular momentum, the wind does not form at all and thus this instability is absent. Higher the angular momentum, stronger is the shear instability between the incoming and outgoing components. When the amplitude of the fastest growing mode becomes non-linear, instabilities in the upper and lower halves join and push the entire disk on one side or the other. This is what we believe to be the cause of the vertical motion. Clearly, this requires a thorough study. The simulations we carried out are inviscid and thus the oscillations are more violent as the angular momentum is not transported away. Similarly, we have not included radiative cooling, because of which, the flow is very hot inside CENBOL. It has already been suggested that the shock oscillation contributes to quasi-periodic oscillations (QPOs) of black hole candidates. We believe that the vertical oscillation of CENBOL may contribute to the variability classes as seen in objects such as GRS 1915+105. The vertical erratic movement of the CENBOL could be responsible for the varying Comptonized component as seen by a specific observer and this may give rise to a large number of variability classes even in so-called hard states, soft states or intermediate states. Giri \& Chakrabarti (2013) and Giri et al. (2015) have shown that an injected sub-Keplerian flow can distribute the angular momentum and produce and sustain a Keplerian disk when viscosity is higher than a critical value, giving rise to the CT95 configuration of two component advective flow. However, this conclusion was drawn using simulations with the upper quadrant only. The conclusions drawn in the present paper with serve as the basis of of the next work with viscosity and radiative cooling. Most importantly, it would be clear if the resulting standard disk also exhibit such an oscillation. This will be discussed elsewhere.

1607.06453

1607

1607.01045_arXiv.txt

{Long gamma-ray bursts (LGRBs) and superluminous supernovae (SLSNe) are both explosive transients with very massive progenitor stars. Clues about the nature of the progenitors can be found by investigating environments in which such transients occur. While studies of LGRB host galaxies have a long history, dedicated observational campaigns have only recently resulted in a high enough number of photometrically and spectroscopically observed SLSN hosts to allow statistically significant analysis of their properties. In this paper we make a comparison of the host galaxies of hydrogen-poor (H-poor) SLSNe and the {\it Swift}/BAT6 sample of LGRBs. In contrast to previous studies we use a complete sample of LGRBs and we address a special attention to the comparison methodology and the selection of SLSN sample whose data have been compiled from the available literature. At intermediate redshifts ($0.3 < z < 0.7$) the two classes of transients select galaxies whose properties (stellar mass, luminosity, star-formation rate, specific star-formation rate and metallicity) do not differ on average significantly. Moreover, the host galaxies of both classes of objects follow the fundamental metallicity relation and the fundamental plane of metallicity. In contrast to previous studies we show that at intermediate redshifts the emission line equivalent widths of the two populations are essentially the same and that the previous claims regarding the higher fraction of SLSN hosts among the extreme emission line galaxies with respect to LGRBs are mostly due to a larger fraction of strong-line emitters among SLSN hosts at $z < 0.3$, where samples of LGRB hosts are small and poorly defined.}

Distinctly more luminous than ordinary supernovae, the recently established class of superluminous supernovae \citep[SLSN;][]{Quimby2011} is associated with the deaths of very massive stars \citep[e.g.][]{GalYam2009} and is being recognized as a new class of cosmic beacons that pinpoints distant star-forming galaxies \citep[][hereafter L14, L15, A16 and P16]{Lunnan2014,Leloudas2015,Angus2016,Perley2016d} and lights up their environment \citep{Berger2012,Vreeswijk2014}. Due to their extreme luminosities, together with peculiar spectral and light curve properties, understanding the nature of the progenitors and emission mechanisms involved in SLSNe has proved to be rather challenging. SLSNe come in many flavours, which can be grouped into two subclasses based on the presence of hydrogen in their spectra \citep{GalYam2012}: hydrogen-rich SLSN-II and hydrogen-poor (H-poor) SLSN-I. The extreme luminosities of the former are commonly explained as an interaction of supernova ejecta with a dense interstellar medium \citep[e.g.][but see \citealt{Inserra2016}]{Moriya2013}. The physical nature of the SLSN-I type however remains debated. A subclass of slowly evolving SLSN-I was proposed to be powered by radioactive decay of \element[][56]{Ni} \citep[][]{GalYam2009,GalYam2012}. The amount of \element[][56]{Ni} required in this case should be large, but could be produced by a pair-instability SN \citep{Barkat1967}. Such SN is expected to arise in the case of a very massive star ($\sim 140-260$ M$_{\odot}$), in which the cascading conversion of star-supporting photons to electron-positron pairs is followed by a rapid contraction and thermonuclear explosion. On the other hand, SLSN-I emission could be powered by a central engine through the reheating of the SN ejecta by an accretion onto a central black hole \citep{Dexter2013} or spin-down of a rapidly-rotating neutron star with strong magnetic fields (i.e. magnetar model; \citealt{Woosley2010,Dessart2012,Inserra2013,Nicholl2013,Nicholl2015}). In addition to studying SLSN light curves and spectra, clues about the progenitors of SLSN can be found by investigating the environments in which they occur. SLSN host galaxies are often characterized by strong nebular emission lines (L14; L15; P16). In cases where the host is experiencing a recent and young starburst, the equivalent widths of the emission lines can give a meaningful constraint on the age of the stellar population and therefore the age and mass of a progenitor star \citep{Thoene2015}. Properties of the environment can also provide clues to distinguish between more or less probable progenitor models. For example, progenitor stars are expected to be of low metallicity in order to produce the rapidly rotating neutron star \citep{Woosley2006,Yoon2006} or pair-instability SN \citep{Langer2007} and therefore their host environment is expected to be metal-poor. Studies performed so far have shown that the H-poor SLSNe are typically found in host galaxies of low luminosity, low stellar mass (M$_{\star}$), low star formation rates (SFRs) and low nebular metallicities \citep[][L14; L15; A16; P16]{Neill2011}. On the other hand, host galaxies of SLSN-II type are found in galaxies spanning a larger range in luminosity, mass and SFR. \begin{table*} \small \renewcommand{\arraystretch}{1.3} \begin{center} \begin{tabular}{lccccccr} \hline \hline SLSN & redshift & M$_{\rm B}$ & $\log$ M$_{\star}$ & SFR & 12 + $\log \left( \frac{\rm O}{\rm H}\right)$ & References\\ & & (mag) & (M$_{\odot}$) & (M$_{\odot}$yr$^{-1}$) & M08 & \\ \hline MLS12110 & 0.303 & -19.39 & 9.25$_{-0.06}^{+0.04}$ & 0.73 $\pm$ 0.02 & 8.54$_{-0.05}^{+0.05}$ & 1, 2\\ PTF12mxx & 0.33 & -16.9 & 8.16$_{-0.24}^{+0.61}$ & 0.11$_{-0.02}^{0.04}$ & 8.69$_{-0.46}^{+0.23}$& 4, 4\\ PTF09cwl & 0.349 & -15.7 & 7.85$_{-0.33}^{+0.23}$& $<0.06$ & - & 4,4\\ PTF10bjp & 0.359 & -18.8 & 9.21$_{-0.17}^{+0.15}$ & 0.93$_{-0.25}^{+0.28}$ & 8.24$_{-0.23}^{+0.13}$ & 4, 4\\ SN2006oz & 0.396 & -16.96 & 8.67$_{-0.04}^{+0.11}$ & 0.13 $\pm$ 0.11 & 8.46$_{-0.13}^{+0.08}$ & 1, 3\\ PTF10vqv & 0.45 & -18.4 & 8.08$_{-0.15}^{+0.92}$ &1.36$_{-0.36}^{+0.41}$ & 8.51$_{-0.08}^{+0.06}$ & 4, 4\\ PTF09atu & 0.501 & -16.3& 8.35$_{-0.62}^{+0.33}$& $<0.11$ & - & 4, 4\\ PS1-12bqf & 0.522 & -20.28 & 9.45$_{-0.10}^{+0.13}$ & 1.05 $\pm$ 0.55 & 8.82$_{-0.18}^{+0.20}$ & 2, 2\\ PS1-11ap & 0.524 & -18.83 & 8.48$_{-0.12}^{+0.15}$ & 0.30 $\pm$ 0.18 & 8.42$_{-0.12}^{+0.10}$ & 2, 2\\ PS1-10bzj & 0.650 & -17.90 & 7.21$_{-0.05}^{+0.11}$ & 5.85 $\pm$ 2.10 & 7.76$_{-0.12}^{+0.15}$ & 2, 2\\ PS1-12zn & 0.674 & -18.75 & 8.34$_{-0.32}^{+0.24}$ & 4.50 $\pm$ 2.00 & 8.40$_{-0.07}^{+0.07}$ & 2, 2\\ \hline \hline \end{tabular} \end{center} \caption{Properties of H-poor SLSN sample used in the paper for cumulative plots. All objects have available near-infrared photometric observations and therefore reliably measured stellar masses. Reported are SLSN designation, redshift, absolute {\it B}-band magnitude, stellar mass, star formation rate and nebular metallicity (in the M08 metallicity calibration). The last column reports the references from which the stellar masses and spectra were obtained, respectively. Masses and $M_{\rm B}$ were adopted directly from the referenced works, while SFR and metallicities were measured as described in Section 2.\newline References: (1) \citet{Angus2016} (2) \citet{Lunnan2014} (3) \citet{Leloudas2015} (4) \citet{Perley2016d}. } \label{tab1} \end{table*} To put the results into perspective, SLSN hosts have been compared to the host galaxies of other types of explosive transients with massive progenitors. Of particular interest is a comparison with long duration gamma-ray bursts \citep[LGRBs; e.g.][]{Kumar2015}, whose optical afterglow luminosity shortly after the explosion can vastly exceed that of SLSNe \citep[e.g.][]{Bloom2009}, and for which the clues of their progenitors are likewise being sought in their host galaxy environment \citep[e.g.][]{LeFloch2003,Savaglio2009,Levesque2010,Graham2013,Kruhler2015,Vergani2015,Perley2015b,Japelj2016}. In addition to the collapsar scenario \citep[e.g.][]{Woosley1993} the magnetar has also been invoked as a possible central engine of LGRBs and their accompanying supernovae \citep{Usov1992,Metzger2015,Greiner2015,Cano2016}. If the central engine for the two transients is indeed the same, it is expected that SLSN and LGRB are preferentially found in the same environments (though the opposite is not necessarily true). Comparing the host galaxies of SLSN-I and LGRBs, L14 finds them to be similar in M$_{\star}$, SFR, specific SFR (SFR divided by stellar mass) and metallicity, especially if excluding the events at the lowest redshifts. On the contrary, spectroscopic study of L15 and photometric study of A16 conclude that the LGRB hosts are found in less extreme environments (e.g. higher M$_{\star}$ and SFRs and lower emission line EWs) than SLSN-I. The reason for different conclusions could at least partly be attributed to the methodology and biased samples of both SLSNe and LGRBs used in the comparison. We therefore aim to take stock of this issue by comparing the properties of the host galaxies of the {\it Swift}/BAT6 complete sample of LGRBs \citep{Salvaterra2012} and a sample of SLSN host galaxies carefully selected from the literature. We discuss the effects the selection of the SLSN sample has on our conclusions and emphasise the importance of a redshift interval assumed in the comparison of the two galaxy populations. Because of the limited statistics for the H-rich class of SLSNe (SLSN-II), our discussion is limited to the host galaxies of H-poor SLSNe. All errors are reported at 1$\sigma$ confidence unless stated otherwise. We use a standard cosmology \citep{Planck2014}: $\Omega_{\rm m} = 0.315$, $\Omega_{\Lambda} = 0.685$, and $H_{0} = 67.3$ km s$^{-1}$ Mpc$^{-1}$. All quantities are computed with respect to the Chabrier initial mass function \citep{Chabrier2003}. \begin{figure*}[] \centering \includegraphics[scale=0.58]{plots/slsn_mass_mag.pdf} \caption{Cumulative distributions of {\it (a)} stellar mass and {\it (b)} absolute {\it B}-band magnitude of SLSN and LGRB samples in the $0.3 < z < 0.7$ redshift range. For stellar masses the distributions were obtained via MC simulation taking into account the errors of individual measurements.} \label{fig1} \end{figure*} \begin{figure*}[] \centering \includegraphics[width=\textwidth]{plots/slsn_sfr_ssfr_Z.pdf} \caption{Cumulative distributions of {\it (a)} star formation rate, {\it (b)} specific SFR and {\it (c)} metallicities of SLSN and LGRB samples in the $0.3 < z < 0.7$ redshift range.} \label{fig2} \end{figure*}

We compared the properties of the host galaxies of SLSNe to those of LGRBs with the aim of looking for similarities and possible differences of the two populations, to try to understand the different results found in the literature on this topic and with the final goal to shed some light on the progenitors of these two phenomena. In contrast to previous studies we used the host galaxies of a complete sample of LGRB as comparison. Furthermore, we performed the comparison in a reduced redshift range, and carefully selected the published stellar masses of SLSN host galaxies. Our results indicate that, at least within $0.3<z<0.7$, the properties of SLSN and LGRB host galaxies are similar, in terms of their $M_{\it B}$, stellar masses, SFR, sSFR and metallicities. We also show that the host galaxies of SLSN follow the FMR and FPZ relations as field star-forming galaxies. The discrepancies of the results of previous studies are due to the different (biased) comparison samples of LGRB host galaxies used, to the different stellar mass values used among all those studies and to the comparison made over a wide/different redshift range for the two classes of explosions. We do not find evidence of the excess of high-EW SLSN hosts, as found by studies who focused on low-$z$ SLSN samples \citep{Leloudas2015}. We stress that to improve these results larger samples are needed, e.g. more high-$z$ SLSN and more low-$z$ LGRB host galaxies as well as a complete sample of SLSNe.

1607.01045

1607

1607.08487_arXiv.txt

A new multidimensional simulation code for relativistic two-fluid electrodynamics (RTFED) is described. The basic equations consist of the full set of Maxwell's equations coupled with relativistic hydrodynamic equations for separate two charged fluids, representing the dynamics of either an electron-positron or an electron-proton plasma. It can be recognized as an extension of conventional relativistic magnetohydrodynamics (RMHD). Finite resistivity may be introduced as a friction between the two species, which reduces to resistive RMHD in the long wavelength limit without suffering from a singularity at infinite conductivity. A numerical scheme based on HLL (Harten-Lax-Van Leer) Riemann solver is proposed that exactly preserves the two divergence constraints for Maxwell's equations simultaneously. Several benchmark problems demonstrate that it is capable of describing RMHD shocks/discontinuities at long wavelength limit, as well as dispersive characteristics due to the two-fluid effect appearing at small scales. This shows that the RTFED model is a promising tool for high energy astrophysics application.

\label{intro} There is growing evidence that the magnetic field plays a crucial role in many astrophysical phenomena. In particular, it is now widely accepted that relativistic outflows from compact objects involve magnetic fields. Examples include pulsar winds, and jets from active galactic nuclei and gamma-ray bursts. The magnetic field strength contained in the relativistic outflows may be so strong that the dominant fraction of its energy flux is carried away in the form of Poynting flux. Relativistic magnetohydrodynamics (RMHD) provides the basic framework for understanding the dynamics of such highly magnetized relativistic plasmas. More importantly, such a Poynting-dominated outflow seems to be converted into a matter-dominated state during its outward propagation. This indicates that there must be an extremely efficient dissipation mechanism, converting the magnetic field energy to particle kinetic energy. The best known example is probably the $\sigma$-problem of the pulsar wind in the Crab nebula \citep[e.g.,][]{1974MNRAS.167....1R,1984ApJ...283..694K}. Because the relativistic wind itself is driven by the rapid rotation of a strongly magnetized neutron star, it is likely to be in a Poynting-dominated regime at least initially launched from the star. Observation of synchrotron emission from the nebular region downstream of the termination shock, however, indicates clearly that the energy density is dominated by relativistic electron-positron pairs. Similar situations may occur in relativistic jets from black holes \citep[e.g.,][]{1994MNRAS.270..480T,2002A&A...391.1141D}. Although it is plausible that the magnetic field plays the dominant role in accelerating the relativistic jet \citep[e.g.,][]{1999ApJ...522..727K,2004ApJ...606..395M, 2005ApJ...625...60N,2006MNRAS.368.1561M,2007MNRAS.380...51K,2009MNRAS.394.1182K, 2009ApJ...699.1789T}, particle energization must be associated with electromagnetic energy dissipation to explain the observed high energy emission from relativistic particles. It is thus natural to invoke magnetic reconnection as a possible mechanism to explain the inferred dissipation in high energy astrophysical objects \citep{1990ApJ...349..538C,1994ApJ...431..397M,2001ApJ...547..437L, 2003ApJ...591..366K}. Introducing finite resistivity is necessary to take into account the effect of magnetic reconnection, which motivates the development of resistive RMHD simulation codes and application to relativistic magnetic reconnection \citep[e.g.,][]{2006ApJ...647L.123W,2007MNRAS.374..415K,2010ApJ...716L.214Z, 2011ApJ...739L..53T,2013ApJ...775...50T}. On the other hand, Ohm's law in the relativistic regime is not understood very well. Most of current resistive RMHD simulations have adopted the simplest kind of Ohm's law without temporal evolution terms. From numerical point of view, this form of Ohm's law has difficulty at the infinite conductivity limit. Namely, there is an inherent singularity at infinite conductivity (or zero resistivity), because the numerical procedure involves division by resistivity. Intuitively, this sounds odd, and a different form of resistivity may be preferred both physically and numerically. Another possible issue in an ideal RMHD is the assumption of a dense plasma, such that any electric fields in the plasma rest frame are immediately screened out. In a strongly magnetized regime, however, the number of charges may not be enough to cancel the electric field. In other words, when the magnetic field gradient requires, via Ampere's law, the current beyond the limit that can be provided by the plasma, it must be compensated by the displacement current, and there appears a finite (oscillatory) electric field in the rest frame of the fluid. In terms of linear dispersion relation, these waves correspond to high-frequency eigenmodes of a two-fluid plasma. The electron plasma and/or cyclotron frequency is a typical threshold frequency, above which the displacement current plays the dominant role. Such high-frequency waves with phase speeds greater than the speed of light are called superluminal waves, whereas all the RMHD waves have subluminal phase speeds. Equivalently, RMHD is valid only when $B > E$ is satisfied, i.e., the magnetic field is greater than the local electric field. In high-energy astrophysics application, this condition may not necessarily be satisfied. There has been a renewed interest in the possible importance of the electric-field-dominated regime in the dynamics of a strongly magnetized plasma \citep{2013ApJ...770...18A,2014MNRAS.442.2855T}, which is unfortunately completely beyond the limit of RMHD. These issues have motivated search for a better numerical framework other than RMHD. Fully kinetic Particle-In-Cell (PIC) simulation is obviously the most fundamental model, but it will not be an alternative to RMHD for realistic modeling of macroscopic astrophysical phenomena, at least in the foreseeable future. We believe that a relativistic two-fluid model coupled with the full set of Maxwell's equations \citep[e.g.,][]{2009ApJ...696.1385Z} may possibly be employed for this purpose. In this model, the relativistic hydrodynamic (RHD) equations with the Lorentz force as an external force are solved separately for electron and proton (or positron) fluids, respectively. The dynamics is coupled through the electromagnetic field, which evolves according to Maxwell's equations with the conduction current given by the moments of the two fluids. It is essentially a fluid counterpart of PIC simulation, which thus greatly reduces the computational requirement. As a compromise, it ignores the kinetic effect such as collisionless wave-particle resonances and finite Larmor radius effect, whereas dispersive characteristics arising from the finite inertia of each species are correctly retained. Because the model does not assume any relationships between the electric and magnetic fields, it is possible to investigate the electric-field-dominated regime as well. Although it is sometimes referred as relativistic two-fluid MHD, we prefer to call the model as the relativistic two-fluid electrodynamics (RTFED) to correctly represent its capability of describing the non-MHD regime $E > B$. It can thus bridge the gap between the RMHD and electric-field-dominated regimes. Although application of the RTFED model to astrophysical problems so far has been very limited at present \citep{2009ApJ...696.1385Z, 2009ApJ...705..907Z,2013ApJ...770...18A,2014MNRAS.438..704B,2016MNRAS.458.1939B}, we believe that it has the potential for more widespread use in the astrophysical community. Part of the reason is that numerical methods that can be used for the present system of equations have not adequately been explored. Although many aspects of the equations are in common with, RHD, RMHD, as well as the non-relativistic counterpart of two-fluid model \citep[e.g.,][]{2011PhPl...18i2113S}, it is important to investigate the applicability of existing numerical technologies to this particular set of equations. We here present the application of the single state HLL (Harten-Lax-Van Leer) approximate Riemann solver to the RTFED equations. Because the electric field is also an independent variable that evolves in time according to Ampere's law in this model, one has to be careful about the divergence constraints of both the electric and magnetic fields in multidimensional simulations. We have implemented a variant of the HLL-UCT (Upwind Constrained Transport) scheme \citep{2004JCoPh.195...17L}, which was originally proposed for MHD. It ingeniously combines the CT scheme \citep{1988ApJ...332..659E} and a two-dimensional (2D) version of the HLL Riemann solver. We demonstrate that, with a careful choice of collocation of physical quantities on a computational mesh, the HLL-UCT works quite well for the RTFED equations. We believe that this will be a better approach than the popular generalized Lagrangian multiplier method \citep{2009ApJ...705..907Z,2014MNRAS.438..704B}, in particular for the present set of equations. This is because a finite error in Gauss's law for the electric field may couple with high-frequency Langmuir wave fluctuations, and become a potential source of numerical instability. Our numerical scheme is free this problem. We also present a specific form of friction term between the species that, in the long wavelength limit, reduces to the resistive Ohm's law that has commonly been used in current resistive RMHD simulations. Our newly developed three-dimensional (3D) simulation code does not assume any symmetry between the two fluids, and can be used both for electron-positron or electron-proton plasmas with an arbitrary mass ratio. Simulation results for several benchmark problems are presented to demonstrate the robustness of our numerical algorithm. This paper is organized as follows. The basic equations are given in section 2 with brief summary of the model characteristics. The numerical algorithm is described in section 3, and numerical results for the test problems are shown in section 4. Finally, section 5 gives conclusions of this paper.

\label{conclusion} In this paper, we have discussed the relativistic two-fluid electrodynamics (RTFED) equations, which are an extension of relativistic magnetohydrodynamics (RMHD). The advantage of the RTFED model is obviously its capability for a wider range of applications. In contrast to RMHD, there is no inherent difficulty for dealing with a region where the local electric field is larger than the magnetic field, which may become important for extreme environments in high energy astrophysics. Also, it is easy to implement a finite resistivity without suffering from the singularity at infinite conductivity. The resistivity (or the friction term) introduced in this paper reduces to the one used in current generation resistive RMHD codes in the long wavelength limit. This fact remains valid regardless of the proton-to-electron mass ratio, which makes it possible to investigate the resistive effect in not only a pair plasma but also an electron-proton plasma. A 3D simulation code solving the RTFED equations has been described. The code achieves overall second-order accuracy for smooth profiles. If the grid size is taken to be large compared to the skin depth, the RMHD shocks/discontinuities are captured without appreciable numerical oscillation. Furthermore, dispersive waves arising from the two-fluid effect are correctly described in cases where sufficient resolution is available. The numerical algorithm presented here guarantees that the two divergence constraints for the electromagnetic field are preserved up to machine precision. It is also possible to extend the code to higher order. Indeed, \cite{2016JCoPh.318..169B} have presented up to a fourth-order accurate finite volume implementation for the RTFED equations. They have also invented a novel and more consistent reconstruction scheme of the electric field that satisfies Gauss' law over the entire control volume. Alternatively, one may also adopt a finite difference approach, in which case a higher-order reconstruction/interpolation can be applied in a dimension-by-dimension manner. Note that, even in this case, it is possible to construct a scheme that exactly preserves the divergence constraints \citep[e.g.,][]{2007A&A...473...11D}. To be fair, there is one critical disadvantage in the RTFED equation. Because it includes high-frequency plasma waves as eigenmodes even in the long wavelength limit, the numerical stability inevitably requires a small time step to resolve the plasma frequency. This is the most serious obstacle for the model when application to macroscopic phenomena (in the presence of a dense plasma) is considered. In these situations, the dynamical time scale and the inverse plasma frequency will differ by orders of magnitude. A naive way to resolve the issue is to use an implicit time integration scheme \citep[e.g.,][]{Kumar2012,2016JCoPh.318..169B}. Indeed, because the high-frequency waves in the long wavelength limit are non-propagating, the scheme can be made locally implicit. This is advantageous because it will not require communications with neighboring processors in parallelization on a distributed memory system. It may also be possible to utilize analytic solutions for such high-frequency waves combined with the operator splitting technique if the wave amplitude remains sufficiently small. Another numerical issue, although less important than the one above, is associated with the use of the simplified Riemann solver. Because the maximum phase speed in the RTFED equations is always given by the speed of light, it may introduce excessive numerical dissipation in situations where the RMHD characteristics have only non-relativistic speeds. In a sense, if the dynamical time scale is much less than the light transit time, one may think that the dynamics of the plasma are decoupled from the electromagnetic wave propagation. In principle, taking advantage of the decoupling, it would be possible to construct a more sophisticated Riemann solver that resolves the internal structure of the Riemann fan. This allows one to obtain higher accuracy both in non-relativistic and relativistic regions at the same time. Despite the numerical issues raised above, given the potential of the RTFED equations, it is important to continue investigating of numerical methods to overcome the weak points. This will certainly extend the applicability of the model in the field of high energy astrophysics.

1607.08487

1607

1607.01790_arXiv.txt

We present a new measurement of the Hubble Constant \hc\ and other cosmological parameters based on the joint analysis of three multiply-imaged quasar systems with measured gravitational time delays. First, we measure the time delay of \hequad\ from 13-year light curves obtained as part of the COSMOGRAIL project. Companion papers detail the modeling of the main deflectors and line of sight effects, and how these data are combined to determine the time-delay distance of \hequad. Crucially, the measurements are carried out blindly with respect to cosmological parameters in order to avoid confirmation bias. We then combine the time-delay distance of \hequad\ with previous measurements from systems \blens\ and \rxjlens\ to create a Time Delay Strong Lensing probe (TDSL). In flat \LCDM\ with free matter and energy density, we find \hc$=71.9^{+2.4}_{-3.0}\ {\rm km\, s^{-1}\, Mpc^{-1}}$ and $\Omega_{\Lambda}=0.62^{+0.24}_{-0.35}$. This measurement is completely independent of, and in agreement with, the local distance ladder measurements of \hc. We explore more general cosmological models combining TDSL with other probes, illustrating its power to break degeneracies inherent to other methods. The joint constraints from TDSL and Planck are \hc = $69.2_{-2.2}^{+1.4}\ {\rm km\, s^{-1}\, Mpc^{-1}}$, $\Omega_{\Lambda}=0.70_{-0.01}^{+0.01}$ and $\Ok=0.003_{-0.006}^{+0.004}$ in open \LCDM and \hc$=79.0_{-4.2}^{+4.4}\ {\rm km\, s^{-1}\, Mpc^{-1}}$, $\Ode=0.77_{-0.03}^{+0.02}$ and $w=-1.38_{-0.16}^{+0.14}$ in flat \wCDM. In combination with Planck and Baryon Acoustic Oscillation data, when relaxing the constraints on the numbers of relativistic species we find \nnu= $3.34_{-0.21}^{+0.21}$ in \nnuLCDM and when relaxing the total mass of neutrinos we find \mnu\ $\leq\, 0.182$~eV in \mnuLCDM. Finally, in an open \wCDM in combination with Planck and CMB lensing we find \hc$=77.9_{-4.2}^{+5.0}\ {\rm km\, s^{-1}\, Mpc^{-1}}$, $\Ode=0.77_{-0.03}^{+0.03}$, $\Ok=-0.003_{-0.004}^{+0.004}$ and $w=-1.37_{-0.23}^{+0.18}$.

\label{sec:intro} In the past decade, the Standard Cosmological Model, \LCDM, which assumes the existence of either a cosmological constant or a form of dark energy with equation of state $w = -1$, and large scale structure predominantly composed of Cold Dark Matter, has been firmly established given observations to date \citep[e.g.][]{Hinshaw2013, Planck2015}. From a minimal set of 6 parameters describing $\Lambda$CDM, one can in principle infer the value of other parameters such as the current expansion rate of the Universe, $H_0$. However, such an inference involves strong assumptions about the cosmological model, such as the absence of curvature or a constant equation of state for the dark energy. Conversely, we can relax these assumptions and explore models beyond flat-\LCDM using a wider set of cosmological probes. In this case, the analysis benefits greatly from independent measurements of $H_0$ from observations of distance probes such as the distance ladder or water masers \citep[see e.g.][for a review]{Treu2010, Weinberg2013, Treu2016}. As \citet{Weinberg2013} point out, the Figure of Merit of any stage III or stage IV cosmological survey improves by 40\% if an independent measurement of $H_0$ is available to a precision of 1\%. The ``time-delay distances'' in gravitationally lensed quasar systems offer an opportunity to measure $H_0$ independently of any other cosmological probe. First suggested by \citet{Refsdal1964}, this approach involves measuring the time delays between multiple images of a distant source produced by a foreground lensing object. The time delays depend on the matter distribution in the lens (galaxy), on the overall matter distribution along the line-of-sight and on the cosmological parameters. The time delays are related to the so-called time-delay distance $D_{\Delta t}$ to the lens and the source, which is primarily sensitive to \hc and has a weak dependence on the matter density $\Om$, the dark energy density $\Ode$, the dark energy equation of state, $w$, and on the curvature parameter $\Omega_k$ \citep[e.g.][]{Linder2011, Suyu2010b}. The first critical step for the method to work is the measurement of the time delays from a photometric monitoring campaign to measure the shift in time between the light curves of the lensed images of quasars. Such monitoring campaigns must be long enough, and have good enough temporal sampling, to catch all possible (and usually small) photometric variations in the light curves. This is the goal of the COSMOGRAIL collaboration: the COSmological MOnitoring of GRAvItational Lenses, which has been monitoring about 20 lensed quasars with 1m-class and 2m-class telescopes since 2004 \citep[e.g.][]{Courbin2005, Eigenbrod2006a, Bonvin2016}. The target precision for the time delay measurements is a few percent or better, because the error on the time delays propagates linearly to first order on the cosmological distance measurement. Examples of COSMOGRAIL results include \citet{Courbin2011}, \citet{Tewes2013b}, \citet{Rathnakumar2013}, and \citet{Eulaers2013}. The second critical step is the modeling of the lens galaxy. Indeed, time-delay measurements alone can constrain only a combination of the time-delay distance and the surface density of the lens around the quasar images \citep{Kochanek2002}. Additional constraints on the density profile of the lens are therefore required in order to convert observed time delays into inferences of the time-delay distance. These constraints can be derived from velocity dispersion measurements, and the radial magnification of the extended, lensed arc image of the quasar host galaxy \citep[e.g.][]{Suyu2010b, Suyu2014}. Ideal targets for this purpose are lensed quasars with a prominent host, which offer strong constraints on the density profile slope of the foreground lens. In modeling the lens mass distribution, special care has to be paid to the mass-sheet degeneracy (MSD), and, more generally, the source-position transformation (SPT) \citep[e.g.][]{Falco1985, Wucknitz2002, Schneider2013, Schneider2014, Xu2016, Unruh2016}. These can be seen as degeneracies in the choice of the gravitational lensing potential that leave all the lensing observables invariant except for the modeled time delay, $\Delta t$. In other words, a wrong model of the main lens mass distribution can perfectly fit the observed morphology the lensing system, and yet result in an inaccurate inference of the time-delay distance. Priors and spectroscopic constraints on the dynamics of the main lens therefore play a critical role in avoiding systematic biases. In addition, perturbations to the lens potential by the distribution of mass along the line-of-sight also create degeneracies in the lens modeling. The latter can be mitigated with a measurement of the mass distribution along the line-of-sight, for example by using spectroscopic redshift measurements of the galaxies in the lens environment \citep[e.g.][]{Fassnacht2006, Wong2011}, comparisons between galaxy number counts in the real data and in simulations \citep{Hilbert2007, Hilbert2009, Fassnacht2011, Collett2013, Greene2013, Suyu2013, McCully2016} or using weak-lensing measurements (Tihhonova et al., in prep.) The H0LiCOW collaboration (\hc Lenses in COSMOGRAIL's Wellspring) capitalizes on the efforts of COSMOGRAIL to measure accurate time delays, and on high quality auxiliary data from {\it Hubble Space Telescope} ({\it HST}) and 10-m class ground-based telescopes, to constrain cosmology. The H0LiCOW sample consists of five well-selected targets, each with exquisite time-delay measurements. \blens, monitored in radio band with the VLA \citep{Fassnacht2002}, and \rxjlens, monitored by COSMOGRAIL in the visible \citep{Tewes2013b}, have already shown promising results, with relative precisions on the time-delay distance of 5\% and 6.6\% respectively \citep[][]{Suyu2010b, Suyu2014}. \begin{table*} \caption{Optical monitoring campaigns of \hequad. The sampling is the mean number of days between the observations, not considering the seasonal gaps.} \centering \begin{tabular}{l c c c r r r c r} \hline Telescope & Camera & FoV & Pixel & Period of observation & $\#$obs & Exp.time & median FWHM & Sampling\\ \hline Euler & C2 & 11'$\times$11' & 0.344'' & Jan 2004 - Mar 2010 & 301 & 5$\times$360s &1.37'' & 6 days \\ Euler & ECAM & 14.2'$\times$14.2' & 0.215'' & Sep 2010 - Mar 2016 & 301 & 5$\times$360s &1.39'' & 4 days \\ Mercator & MEROPE & 6.5'$\times$6.5' & 0.190'' & Sep 2004 - Dec 2008 & 104 & 5$\times$360s &1.59'' & 11 days \\ Maidanak & SITE & 8.9'$\times$3.5' & 0.266'' & Oct 2004 - Jul 2006 & 26 & 10$\times$180s &1.31'' & 16 days \\ Maidanak & SI & 18.1'$\times$18.1' & 0.266'' & Aug 2006 - Jan 2007 & 8 & 6$\times$300s &1.31'' & 16 days \\ SMARTS & ANDICAM & 10'$\times$10' & 0.300'' & Aug 2003 - Apr 2005 & 136 & 3$\times$300s &$\leq$1.80'' & 4 days \\ \hline {\bf TOTAL} & - & - & - & Aug 2003 - Mar 2016 & 876 & 394.5h & - & 3.6 days \\ \hline \end{tabular} \label{tab:monitoring} \end{table*} \begin{figure*} \centering \includegraphics[width=0.99\textwidth]{he0435_deepfield.pdf} \caption{Part of the field of view of EulerCAM installed on the Swiss 1.2m telescope around the quasar \hequad. This image is a combination of 100 exposures of 360s each, for a total exposure time of 10 hours. The stars used to build a PSF model for each EulerCAM exposure are circled and labeled PSF1 to PSF7 in red, and the stars used for the photometric calibrations are circled and labeled N1 to N8 in green. The insert in the bottom left shows the single, 360s exposure of the lens, for reference. Note that photometric and spectroscopic redshifts are available for many galaxies in the field of view (see \PaperII and \PaperIII for details). } \label{fig:he0435_deepfield} \end{figure*} This paper is part of the H0LiCOW series, focusing on the quadruple lensed quasar \hequad\ ($\alpha$(2000):~04h~38m~14.9s; $\delta$(2000):~-12$^\circ$17\arcmin14\farcs4) \citep{Wisotzki2000, Wisotzki2002} discovered during the Hamburg/ESO Survey (HES) for bright quasars in the Southern Hemisphere. The source redshift has been measured by \citet{Sluse2012} as $\zs = 1.693$, and the redshift of the lens has been measured by \citet{Morgan2005} and \citet{Eigenbrod2006b} as $\zd = 0.4546 \pm 0.0002$. The lens lies in a group of galaxies of at least 12 members. A first measurement of the time delay for \hequad\ was presented in \citet{Courbin2011}. In this work, we present a significant improvement of the time delay measurement, with twice as long light curves as in \citet{Courbin2011}. The other H0LiCOW papers include an overview of the project (Suyu et al., submitted; hereafter \PaperI), a spectroscopic survey of the field of \hequad\ and a characterization of the groups along the line-of-sight (Sluse et al., submitted; hereafter \PaperII), a photometric survey of the field of \hequad\ with an estimate of the effect of the external line-of-sight structure (Rusu et al., submitted; hereafter \PaperIII), and a detailed modeling of the lens and the inference of the time-delay distance along with cosmological results for \hequad\ (Wong et al., submitted; hereafter \PaperIV). In the present paper we combine the results for \hequad\ with those from the other two lensed quasars already published, and with other cosmological datasets \citep{Bennett2013, Hinshaw2013, Planck2015}. This paper is organized as follows. Section~\ref{sec:data} presents the COSMOGRAIL optical monitoring data and its reduction process. Section~\ref{sec:timedelay} presents the time-delay measurements and related uncertainties. Section~\ref{sec:modeling} summarizes the main steps of the field-of-view analysis detailed in \PaperII and \PaperIII and the lens modeling detailed in \PaperIV that lead to the time-delay distance determination. Section~\ref{sec:cosmography} combines the time-delay distance of \hequad\ and other lenses, and with additional cosmological datasets, in order to make the best possible inferences of cosmological parameters. Finally, Section~\ref{sec:conclusion} presents our conclusions and future prospects in the light of these results.

\label{sec:conclusion} Using multiple telescopes in the Southern and Northern hemispheres, we have monitored the quadruple imaged strong gravitational lens \hequad\ for 13 years with an average cadence of one observing epoch every 3.6 days. We analyse the imaging data using the MCS deconvolution algorithm \citep{Magain1998} on a total of 876 observing epochs to produce the light curves of the four lensed images, with an rms photometric precision of 10 mmag on the brightest quasar image. We measured the time delays between each pair of lensed images using the free-knot spline technique and the regression difference technique from the {\tt PyCS} package \citep{Tewes2013a}. Our uncertainty estimation involves the generation of synthetic light curves that closely mimic the intrinsic and extrinsic features of the real data. To test the robustness of our measurements, we vary parameters such as the number of knots in the splines, the initial parameters used for the deconvolution photometry, and the length of the considered light curves. The two curve shifting techniques agree well with each other both on the point estimation of the delays and on the estimated uncertainty. The smallest relative uncertainty, of 6.5\%, is obtained for the A-D pair of images. For this pair involving image A, our present measurement is twice as precise as the earlier result by \citet{Courbin2011}. In \PaperIV, we use our new COSMOGRAIL time delays for \hequad\ to compute its time-delay distance. Very importantly, this is done in a blind way with respect to the inference of cosmological parameters. In this paper, we combine the time-delay distance likelihoods from \hequad\ with the published ones from \blens\ and \rxjlens\ to create a Time Delays Strong Lensing (TDSL) probe. We also combine the latter with other cosmological probes such as WMAP, Planck, BAO and JLA to constrain cosmological parameters for a large range of extended cosmological models. Our main conclusions are as follows: \begin{itemize} \item TDSL alone is weakly sensitive to the matter density, $\Om$, curvature, $\Ok$ and dark energy density $\Ode$ and equation of state $w$. Its primary sensitivity to \hc allows us to break degeneracies of CMB probes in extended cosmological models. \item In a flat \LCDM cosmology with uniform priors on \hc and $\Om$, TDSL alone yields \hc$=71.9_{-3.0}^{+2.4}\ {\rm km\, s^{-1}\,Mpc^{-1}}$. When enforcing $\Om$=0.32 from the most recent Planck results, we find \hc=$72.8\pm2.4\ {\rm km\, s^{-1}\,Mpc^{-1}}$. These results are in excellent agreement with the most recent measurements using the distance ladder, but are in tension with the CMB measurements from Planck. \item In a non-flat \LCDM cosmology, we find, using TDSL and Planck, \hc=$69.2_{-2.2}^{+1.4}\ {\rm km\, s^{-1}\,Mpc^{-1}}$ and $\Ok=0.003_{-0.006}^{+0.004}$, in agreement with a flat universe. \item In a flat \wCDM cosmology in combination with Planck, we find a 2.3$\sigma$ tension from a cosmological constant in favor of a phantom form of dark energy. Our joint constraints in this cosmology are \hc$=79.0_{-4.2}^{+4.4}\ {\rm km\, s^{-1}\,Mpc^{-1}}$, $\Ode=0.77_{-0.03}^{+0.02}$ and $w=-1.38_{-0.16}^{+0.14}$. \item In a flat \mnuLCDM cosmology, in combination with Planck and BAO we tighten the constraints on the maximum mass of neutrinos to \mnu$\leq0.182$~eV, while removing the tension in \hc. \item In a flat \nnuLCDM cosmology, in combination with Planck and BAO we find \nnu$=3.34\pm0.21$, i.e. 1.3$\sigma$ higher than the standard cosmological value. This mild tension remains when the constraints on both \mnu and \nnu are relaxed. \item In a \owCDM cosmology, in combination with Planck and CMBL, we find \hc$=77.9_{-4.2}^{+5.0}\ {\rm km\, s^{-1}\,Mpc^{-1}}$, $\Ode=0.77_{-0.03}^{+0.03}$, $\Ok=-0.003_{-0.004}^{+0.004}$ and $w=-1.37_{-0.23}^{+0.18}$. Similarly to the \oLCDM and \wCDM cosmologies, we are in good agreement with a flat universe and in tension with a cosmological constant, respectively. \end{itemize} We emphasize that despite reporting parameter constraints for a large variety of cosmological models beyond \LCDM, we choose not to comment on whether a particular model is favored over the others. Such an exercise would require a well motivated choice of priors for these models, which is not within the scope of this work.\\ The combined strengths of our H0LiCOW lens modeling and COSMOGRAIL monitoring indicate that quasar time-delay cosmography is now a mature field, producing precise and accurate inferences of cosmological parameters, that are independent of any other cosmological probe. Still, our results can be improved in at least four ways: \begin{enumerate} \item Continuing to enlarge the sample. Two more objects with excellent time-delay measurements as well as high-resolution imaging and spectroscopic data remain to be analysed in the H0LiCOW project (see \PaperI). When completed, H0LiCOW is expected to provide a measurement of \hc to better than 3.5\% in a non-flat \LCDM universe with flat priors on $\Om$ and $\OL$. Data of quality comparable to those obtained for H0LiCOW are in the process of being obtained for another four systems with measured time delays from COSMOGRAIL (HST-GO-14254; PI: Treu). Meanwhile, current and planned wide field imaging surveys such as DES, KiDS, HSC, LSST, Euclid and WFIRST, should discover hundreds of new gravitational lens systems suitable for time-delay cosmography \citep{O+M10}. For example, the dedicated search in the Dark Energy Survey STRIDES\footnote{strides.astro.ucla.edu} has already confirmed two new lenses from the Year1 data \citep{Agnello2016}. \item Improve the lens modeling accuracy. The tests carried out in our current (\PaperIV) and past work \citep{Suyu2014}, the good internal agreement between the three measured systems (Section \ref{sec:ubgcosmo}), and independent analysis based on completely independent codes \citep{Birrer2016}, show that our lens models are sufficiently complex given the currently available data. However, as the number of systems being analysed grows, random uncertainties in the cosmological parameters will fall, and residual systematic uncertainties related to degeneracies inherent to gravitational lensing will need to be investigated in more detail. Following the work of \citet{Xu2016}, detailed hydro N-body simulations of lensing galaxies in combination with ray-shooting can be used to evaluate the impact of the lensing degeneracies on cosmological results in view of future observations with the JWST or 30-m class ground-based telescopes with adaptive optics, and to drive development of improved lens modeling techniques and assumptions appropriate to the density structures we expect. \item Improve the absolute mass calibration. Spatially resolved 2D kinematics of the lens galaxies, to be obtained either with JWST and with integral field spectrographs mounted on large ground-based telescopes with adaptive optics, should further improve both the precision for each system and our ability to test for residual systematics, including those arising from the mass sheet and source position transformations \citep{Schneider2013, Schneider2014, Unruh2016}. The same data should also allow us to use constraints from the stellar mass or mass profile of lens galaxies as attempted in \citet{Courbin2011} with slit spectroscopy. Alternatively, the mass-sheet degeneracy can be lifted if the absolute luminosity of the source is known \citep{Falco1985}, which is the case for lensed standard candles \citep[see e.g.][that report the first discovery of a lensed type-Ia Supernovae]{Goobar2016}. However, such configurations happens far less often than lensed quasars. \item Measuring time delays with the current photometric precision and time sampling of monitoring data requires long and time-consuming campaigns, and is currently not possible for hundreds of objects. Increasing the monitoring efficiency is possible, by catching extremely small (mmag) and fast (days) variations in the quasar light curves. Such data can be obtained with daily observations with 2-m class telescopes in good seeing conditions, a project that will be implemented in the context of the extended COSMOGRAIL program (eCOSMOGRAIL; Courbin et al. 2016, in prep) to measure quasar time delays in only 1 or 2 observing seasons. Furthermore, in the long run, LSST should be able to provide sufficiently accurate time delays for hundreds of systems from the survey data itself \citep{Liao2015}, and enable sub-percent precision on \hc in the next decade \citep{Treu2016}. \end{enumerate}

1607.01790

1607

1607.01267_arXiv.txt

We present a multiwavelength investigation of star formation activity towards the southern \hii~regions associated with IRAS~17160--3707, located at a distance of 6.2~kpc with a bolometric luminosity of $8.3\times10^5$~L$_\odot$. The ionised gas distribution and dust clumps in the parental molecular cloud are examined in detail using measurements at infrared, submillimeter and radio wavelengths. The radio continuum images at 1280 and 610~MHz obtained using Giant Metrewave Radio Telescope reveal the presence of multiple compact sources as well as nebulous emission. At submillimeter wavelengths, we identify seven dust clumps and estimate their physical properties like temperature: $24-30$~K, mass: $300 - 4800$~M$_\odot$ and luminosity: $9 - 317\times10^2$~L$_\odot$ using modified blackbody fits to the spectral energy distributions between 70 and 870~$\mu$m. We find 24 young stellar objects in the mid-infrared, with few of them coincident with the compact radio sources. The spectral energy distributions of young stellar objects have been fitted by the Robitaille models and the results indicate that those having radio compact sources as counterparts host massive objects in early evolutionary stages with best fit age $\le 0.2$~Myr. We compare the relative evolutionary stages of clumps using various signposts such as masers, ionised gas, presence of young stellar objects and infrared nebulosity and find six massive star forming clumps and one quiescent clump. Of the former, five are in a relatively advanced stage and one in an earlier stage.

\hii~regions are the first reliable pointers announcing the formation of massive O and B stars embedded in their natal molecular clouds. The hypercompact (size $<0.01$~pc) and ultracompact (size $<0.1$~pc) \hii~region phases with large electron densities ($n_e\ge 10^4$~cm$^{-3}$) are indicative of the early evolutionary stages comprising young massive stars in the accretion phase, and these young \hii~regions evolve with time into compact and classical \hii~regions \citep{1999PASP..111.1049G,{2007prpl.conf..181H}}. These young \hii~regions are bright in radio and infrared and possess a wide range of morphologies such as cometary, core-halo, shell, bipolar, and irregular \citep{{2002ARA&A..40...27C},{wood1989morphologies}}. While the morphology of a \hii~region provides clues to the density of the surrounding interstellar medium, it also depends on the scale of the region observed: a large scale morphology could lead to another at alternate angular resolution \citep{{wood1989morphologies}, {2004ApJ...605..285S}}. Ultracompact \hii~regions are usually found close to other ultracompact and compact \hii~regions within complexes \citep{{1993ApJ...412..684M},{2001ApJ...549..979K},{sanchez2013deciphering}}. \par The host massive stars (mass greater than $\sim$8~M$_\odot$) remain in an embedded phase that lasts about 15\% of their lifetime before evolving into the optically visible main-sequence phase \citep{{2002ARA&A..40...27C}, {2007ARA&A..45..481Z}}. These initial phases remain observationally elusive as massive stars are rare, evolve fast compared to their lower mass counterparts, and are at relatively large distances ($> 1$~kpc) from us, apart from being born in a clustered environment. The circumstellar dust in the vicinity of a \hii~region absorbs nearly all the stellar radiation and re-emit in the mid and far-infrared. The relatively cooler dust in the envelope absorbs infrared radiation and shows thermal emission in the submillimeter regime, that is optically thin. The far-infrared and submillimeter maps of such regions therefore show warm and cold dust clumps as well as filamentary emission, that can be used as a potential tool for studying the physical properties of the associated molecular clouds \citep[e.g.][]{{2011A&A...533A..94H}, {2012A&A...542A..10A}, {2013A&A...554A..42R}}. By comparing the emission from radio (ionised gas) as well as infrared and submillimeter (warm and cold dust), various stages of evolutionary sequence of massive star forming clumps in molecular clouds can be examined in detail \citep{ {2013A&A...556A..16G},{2010ApJ...721..222B}, {2013A&A...550A..21S}}. These studies highlight the importance of multiwavelength investigation of star forming regions. \par In this paper, we investigate a southern Galactic massive star forming region IRAS~17160--3707. IRAS~17160-3707, also known as G350.10+0.08, harbours a number of compact and ultracompact \hii~regions \citep{2003A&A...407..957M} with a far-infrared (IRAS) luminosity of 3.6$\times$10$^{5}$~L$_{\odot}$ \citep{2002A&A...381..571P}. The near distance to this region ranges from 5.7 to 7.3~kpc \citep{2002A&A...381..571P, 2004A&A...426...97F, 1987A&A...171..261C}. We adopt a distance of 6.2~kpc based on sensitive radio recombination line kinematics that resolve the distance ambiguity using measurement of \mbox{H\,{\sc i}} line absorption towards the region \citep{2006ApJ...653.1226Q}. Being bright in radio, this \hii~complex has been detected in many Galactic plane surveys \citep{{1989ApJS...71..469L},{1990ApJS...74..181Z},{1994ApJS...91..347B},{1998MNRAS.301..640W}, {2006ApJS..165..338Q}}. Among these surveys, \citet{1994ApJS...91..347B} have detected a radio source of size $12''$ in their snapshot observations at 1.4 and 5 GHz. High resolution maps of this region at 4.8 and 8.6~GHz are presented in \citet{2003A&A...407..957M} where they detected five compact \hii~regions. IRAS 17160--3707 also encompasses a number of maser spots that have been detected as a part of maser surveys, such as methanol masers \citep[e.g.][]{1998MNRAS.301..640W}, water masers \citep[e.g.][]{1989A&A...213..339F}, and hydroxyl masers \citep[e.g.][]{2001MNRAS.326..805C}. Emission from Polycyclic Aromatic Hydrocarbons (PAH) as well as silicate absorption were reported in this region by \cite{2002A&A...381..571P}. \par In the present work, we examine the star formation activity in this region using measurements at infrared, submillimeter and radio wavelengths. In particular, our interferometric observations using the Giant Metrewave Radio Telescope (GMRT) enable us to probe the ionised gas at different angular scales: compact objects requiring high resolution as well as the broad diffuse emission. The far-infrared and submillimeter wavebands provide a wide coverage of thermal dust emission, useful for investigating the physical properties of the associated molecular cloud. We attempt to locate the mid-infrared young stellar objects and clumps of gas and dust in this \hii~region complex and ascertain their properties. The details of observations, data reduction and archival data are given in Section 2. In Section 3, we discuss the results. The radio morphology and multiwavelength scenario are presented in Section 4, and in Section 5, we summarize the results of this work.

\par We examined the star formation activity in the \hii~region complex associated with IRAS~17160--3707 using a range of wavelengths spanning from 3.6~$\mu$m to 870~$\mu$m. Based on our analysis, we arrive at the following conclusions: \begin{itemize} \item[(a)] The radio continuum maps show the presence of twelve compact radio sources in the region with ten of them enveloped in diffuse emission. We have estimated the Lyman continuum fluxes of the individual sources. The 1280 - 610 MHz spectral index map reveals the presence of thermal emission towards compact sources and non-thermal emission in certain pockets of the envelope. \item[(b)] We identify seven cold dust clumps from the 870~$\mu$m emission map and fitted modified blackbody to fluxes associated with these clumps giving a dust temperature in the range from 24 to 30~K and column density in the range $1.7 - 7.6 \times$~10$^{22}$~cm$^{-2}$. We have also constructed dust temperature, column density and dust emissivity index maps by carrying out modified blackbody fits to the individual pixels in this region. \item[(c)] The mid-infrared warm dust emission traces the radio morphology and we see emission associated with the compact sources in the mid-infrared bands. We identify 23 YSOs in the region based on mid-infrared color-color diagrams. Five of them are associated with compact radio sources and the radiative transfer modeling of these sources imply that they are relatively young with best fit age $\le 0.2$~Myr. \item[(d)] This region harbors an infrared bubble CS-112 likely to be excited by a group of stars located near the geometric centre. We estimated the age of the bubble as 0.7~Myr based on a simple model of photoionized expansion of a \hii~region. \item[(e)] The dust clumps fall in different classes of evolutionary stages. There are six star-forming clumps and a quiescent clump. The collective picture indicates that this region is relatively young, currently undergoing a high mass star formation bustle. \end{itemize}

1607.01267

1607

1607.04674_arXiv.txt

We present an analysis of the complex gas hydrodynamics in the X-ray luminous galaxy cluster RXJ1347.5-1145 caught in the act of merging with a subcluster to its southeast using a combined~$186$~ks {\it Chandra} exposure,~$2.5$~times greater than previous analyses. The primary cluster hosts a sloshing cold front spiral traced by four surface brightness edges $5 \farcs 85^{+0.04}_{-0.03}$~west, $7 \farcs 10^{+0.07}_{-0.03}$~southeast, $11 \farcs 5^{+1.3}_{-1.2}$~east, and $16 \farcs 7^{+0.3}_{-0.5}$~northeast from the primary central dominant galaxy, suggesting the merger is in the plane of the sky. We measure temperature and density ratios across these edges, confirming they are sloshing cold fronts. We observe the eastern edge of the subcluster infall shock, confirming the observed subcluster is traveling from the southwest to the northeast in a clockwise orbit. We measure a shock density contrast of~$1.38^{+0.16}_{-0.15}$~and infer a Mach number $1.25\pm0.08$ and a shock velocity of $2810^{+210}_{-240}$\,\kms. Temperature and entropy maps show cool, low entropy gas trailing the subcluster in a southwestern tail, consistent with core shredding. Simulations suggest a perturber in the plane of the sky on a clockwise orbit would produce a sloshing spiral winding counterclockwise, opposite to that observed. The most compelling solution to this discrepancy is that the observed southeastern subcluster is on its first passage, shock heating gas during its clockwise infall, while the main cluster's clockwise cold front spiral formed from earlier encounters with a second perturber orbiting counterclockwise.

\label{sec:introduction} The growth of structure across cosmic time provides an important probe of cosmology. The evolution of the galaxy cluster mass function with redshift places significant constraints on the dark energy equation of state, with the greatest sensitivity found at the high mass end \citep{mass_function1, mass_function2}. Since the baryon distribution in massive galaxy clusters is dominated by hot, X-ray emitting gas tracing the cluster dark matter potential, X-ray observations provide a natural tool to identify clusters. If the cluster is relaxed and the gas in hydrostatic equilibrium, these observational measures allow us to calculate the cluster mass. However, in the current hierarchical model for structure formation, massive clusters grow through mergers between galaxy groups and less massive clusters, that formed along overdense filaments in the cosmic web. Thus mergers are expected to be common. Kinetic energy from the merging partner is converted mainly to thermal energy in the form of shocks and turbulence in the cluster gas, profoundly affecting the evolution of the ICM \citep{Fujitab, Fujitaa, Fujitac, zuhone2010, zuhone2011}. Non-hydostatic gas motions induced by the merger (`sloshing') are long lasting on timescales of order gigayears. Their characteristic signatures in X-ray observations, i.e. multiple surface brightness discontinuities (cold fronts) and/or sweeping spiral features in temperature and surface brightness, have been noted in many galaxy groups and clusters \citep[e.g. see review by][]{markevitch2007, ascasibar2006, paterno-mahler, randall_galaxy, machacek_galaxy}. The study of these merger-induced gas motions and entropy exchanges in massive clusters are key to understanding (1) the properties and evolution of the ICM and (2) how the departures from hydrostatic equilibrium affect uncertainties in the cluster mass determinations used to constrain cosmology. \begin{figure*}[t!] \begin{center} \includegraphics[width=\textwidth]{f1.pdf} \caption{Chandra background subtracted, exposure corrected, combined 0.5-2.5 keV full resolution ($1\,\mbox{pixel}=0 \farcs 492\times 0 \farcs 492$) image of RXJ1347 with log scale. The primary cD galaxy is labeled with a black cross and the subcluster cD galaxy is labeled with a white x. Arrows point to edges forming a spiral pattern.} \label{fig:xray} \end{center} \end{figure*} Outstanding questions include: \begin{itemize} \item{What role do merger shocks play in heating the cluster gas \citep{zuhone2010}?} \item{ Does sloshing inhibit or promote cooling in the cluster core \citep{zuhone2010, zuhone2011}?} \item{ What can the morphology of the cold fronts and spirals tell us about the merger history of the cluster, cluster magnetic fields, or microphysics of the gas \citep[see, e.g. review by][]{markevitch2007, roedigerb, roedigera, roedigerc, roedigerd, roedigere}}? \item{ Does gas sloshing amplify magnetic fields and reaccelerate electrons, creating radio mini-halos found to be correlated with cold front edges in some clusters \citep{gitti2002, gitti2004, mazzotta2008, zuhone2013}?} \end{itemize} Progress on these questions will be found by a comparison of deep X-ray observations of relatively nearby ($z \lesssim 0.5$), massive, merging galaxy clusters, where gas temperatures and densities at cold fronts, shocks, and sloshing spirals can be well studied, with high resolution numerical simulations that connect the observed X-ray features with the orbital history of the merger and the microphysical properties of the surrounding gas. In this paper we use new and archival X-ray observations and archival Sunyaev-Zel'dovich effect (SZE) observations from the literature to characterize the cold fronts, shocks, and sloshing in RXJ$1347.5-1145$ (RXJ1347). With a $2 - 10$\,keV X-ray luminosity of $6 \times 10^{45}$\ergs, RXJ1347 is the most X-ray luminous galaxy cluster found in the Rosat All Sky Survey \citep{schindler} and is still one of the more X-ray luminous galaxy clusters known. RXJ1347 is a massive, highly evolved cool core cluster, yet still actively merging with a less massive subcluster. Cool core clusters often show signatures of gas sloshing due to the presence of easily disturbed low entropy gas in their cores \citep{markevitch2007}. This makes RXJ1347 an ideal laboratory to study the on-going evolution of the bright end of the cluster luminosity function. It has been widely studied across many wavebands \citep[see, e.g.][for a review]{johnson2012}. Optical spectroscopic surveys of its galaxy population yield a velocity dispersion and dynamical mass within $r_{ 200}$ for RXJ1347 of $\sigma = 1163 \pm 97$\,\kms and $1.16^{+0.32}_{-0.27} \times 10^{15}\mbox{$\Ms$}$, respectively \citep{lu2010, cohen}. \citet{lu2010} also found that RXJ1347 resides on a large scale filament, separated from a less massive galaxy cluster (RXJ1347-SW) by $\sim 7$\,Mpc in projection and $4000$\,\kms in radial velocity and with an excess of galaxies in between. Although the large velocity difference between the two clusters makes it unlikely that they are interacting, continued merging activity of RXJ1347 with other subclusters embedded within the same large scale filament would be expected in hierarchical cosmological models. The cluster K-band light is dominated by the two brightest cluster galaxies (BCGs). One, at (RA $13^h47^m30.7^s$, Dec $-11^{\circ}45'10 \farcs 1$ ) near the peak of the X-ray surface brightness, is assumed to be the central dominant galaxy of the primary cluster, and the second, at (RA $13^h47^m31.9^s$, Dec $-11^{\circ}45'10 \farcs 9$) lies $18''$ to the east. Given the second galaxy's relative radial velocity difference of only $\sim 100$\kms with respect to the primary cluster BCG, it has been interpreted as the central dominant galaxy of a merging subcluster. Surface mass density maps from weak and strong lensing show a general elongation in the mass concentration in the direction of the second BCG. However, they do not show a clear mass concentration peak there \citep{bradac2008, schmidt}. Early X-ray observations by {\it Chandra} \citep{allen} and {\it XMM-Newton} \citep{gitti2004} found excess X-ray emission~$\sim20''$~southeast of the primary BCG, which we now interpret as emission from hot gas in a merging subcluster. Deeper {\it Chandra} observations have revealed two sloshing cold fronts in the primary cluster as a result of a recent merger \citep{johnson2012}. Sunyaev-Zel'dovich observations have played a critical role in understanding the merging subcluster's gas dynamics. Observations by \citet{komatsu2001} with the Nobeyama 45 m telescope at~150~GHz first showed a significant SZ decrement~$\sim 20''$~southeast of the primary cluster, suggesting RXJ1347 is a disturbed system that has recently undergone a merger \citep{kitayama2004}. Higher resolution observations with the MUSTANG camera on the Green Bank Telescope at~90~GHz indicated the SZ decrement likely corresponds to shock heated gas from the merging subcluster \citep{mason2010, korngut2011}. A higher temperature and characteristic high pressure gradient support the interpretation that the region between the clusters is a shock front \citep{mason2010}. RXJ1347 also hosts a radio mini-halo centered on its primary cluster. While gas sloshing causes electron re-acceleration in the mini-halo, \citep{mazzotta2008}, \citet{ferrari2011} used observations by the Giant Metrewave Radio Telescope at 237 MHz and 614 MHz to find that excess radio emission in the southeast section of the mini-halo is likely caused by a propagating shock front corresponding to the shock-heated gas and SZ decrement previously found. At a redshift of $z=0.451$, RXJ1347 is close enough that the disturbed X-ray morphology resulting from recent mergers can be observed in detail with the high angular resolution of the {\it Chandra} X-ray Observatory. At this high angular resolution, our combined $186$\,ks {\it Chandra} exposure ( $> 2.5$ times deeper than previous {\it Chandra} observations) allows us to identify features not seen before, to better constrain the ICM gas hydrodynamics and, by comparison with simulations, RXJ1347's merger history. In \S\ref{sec:obs} we discuss our reduction of the observational data and general analysis techniques. In \S\ref{sec:qual} we analyze the mean and asymmetric gas properties in the primary cluster. We discuss the properties of the merging subcluster in \S\ref{sec:subcluster}, and we compare our results to simulations in \S\ref{sec:sim} to discern the cluster merging history. For the standard $\Lambda$ dominated cold dark matter cosmology and assuming $H_0 = 70$\kmsmpc, $\Omega_m = 0.3$, $\Omega_\Lambda = 0.7$, the redshift $z=0.451$ for RXJ1347.5-1145 corresponds to a luminosity distance of $2504$\,Mpc and angular scale $1'' = 5.77$\,kpc \citep{cosmocalc}. All WCS coordinates are J2000, and uncertainties are at $90 \%$ CL unless otherwise specified.

\label{sec:conclude} With~$2.5$~times the exposure of previous analyses, we are able to study the gas hydrodynamics of RXJ1347 and its merging subcluster in unprecedented detail. We briefly summarize our key results. \begin{itemize} \item{A series of four cold fronts west, southeast, east, and northeast located at~$5\farcs 85^{+0.04}_{-0.03}$, $7\farcs 10^{+0.07}_{-0.03}$, $11\farcs 5^{+1.3}_{-1.2}$, and $16\farcs 7^{+0.3}_{-0.5}$~from the primary cluster's cD galaxy, respectively, forms a clockwise spiral, suggesting the merger is in the plane of the sky. The west and east cold fronts correspond to those found by \citet{johnson2012}.} \item{We measure a subcluster $0.5-2$\,keV luminosity of $L_x= 1.3 \times 10^{44}$\ergs and infer a lower bound on the total subcluster mass of $3.3 \times10^{14}$~M$_\odot$.} \item{We identify the shock from the supersonic infall of the subcluster and measure a density ratio across the shock of $1.38^{+0.16}_{-0.15}$~, corresponding to a Mach number of $1.25\pm0.08$ and shock velocity of~$2810^{+210}_{-240}$~km~s$^{-1}$. The measured opening angle of the Mach cone of~$\sim50$~deg is consistent with this shock interpretation.} \item{The subcluster's baryonic gas has been stripped from the dark matter peak and leaves a tail behind the subcluster. Asymmetric entropy contours and cooler temperatures extending through the subcluster and tail indicate the subcluster core is being shredded.} \item{We suggest that the most likely explanation for the observed X-ray features is that the southeast subcluster is on its first passage on a clockwise orbit through the cluster, shock heating the cluster gas as it infalls, while the clockwise sloshing spiral observed in the primary cluster was caused by earlier encounters with a second subcluster moving in a counterclockwise orbit. Better simulations and further observations in other wavelengths are needed to search for remnants of this second perturber and test this scenario.} \end{itemize}

1607.04674

1607

1607.08170_arXiv.txt

Motivated by the current search for exomoons, this paper considers the stability of tidal equilibrium for hierarchical three-body systems containing a star, a planet, and a moon. In this treatment, the energy and angular momentum budgets include contributions from the planetary orbit, lunar orbit, stellar spin, planetary spin, and lunar spin. The goal is to determine the optimized energy state of the system subject to the constraint of constant angular momentum. Due to the lack of a closed form solution for the full three-body problem, however, we must use use an approximate description of the orbits. We first consider the Keplerian limit and find that the critical energy states are saddle points, rather than minima, so that these hierarchical systems have no stable tidal equilibrium states. We then generalize the calculation so that the lunar orbit is described by a time-averaged version of the circular restricted three-body problem. In this latter case, the critical energy state is a shallow minimum, so that a tidal equilibrium state exists. In both cases, however, the lunar orbit for the critical point lies outside the boundary (roughly half the Hill radius) where (previous) numerical simulations indicate dynamical instability. These results suggest that star-planet-moon systems have no viable long-term stable states analogous to those found for two-body systems.

\label{sec:intro} A classic dynamical problem is to consider the tidal equilibrium states for self-gravitating systems that include both rotational and orbital motion \citep{darwin1,darwin2}. The conditions required for the existence of such equilibrium states has been determined previously for binary star systems \citep{counselman,hut1980}. With the relatively recent discovery of exoplanets, this problem has received renewed interest \citep{levrard,soko,abtide}. In this contribution, we extend previous treatments to include the presence of a satellite or moon orbiting the secondary (see also \citealt{barnes,scheeres}). We thus consider the existence of tidal equilibrium states for hierarchical triple systems consisting of a star, planet, and moon, where all three bodies have spin angular momentum. Although the discovery of moons in extrasolar planetary systems has not yet been realized, they have generated great interest. Now that the existence of exoplanets is well established, and their populations are being characterized, the next astronomical frontier is to discover moons in other solar systems. These additional bodies are in principle observable through their transit timing variations \citep{kippingttv}, wherein additional bodies in a planetary system change the times at which the planet casts shadows on the host star \citep{agol,holman}. These moons are most readily detected if they have large relative masses and their host planets orbit near the star. Another regime of interest is that of potentially habitable moons, where the orbit of the host planet resides within the habitable zone \citep{kasting} of its star. In this case, a moon orbiting the planet could be potentially habitable provided that it has the proper mass, atmosphere, and other characteristics \citep{kippinghab,heller}. In both regimes, however, moons are susceptible to removal \citep{donnison,weidner,spalding}. In general, systems evolve toward lower energy states, but are required to conserve their total angular momentum (in the absence of external torques). For two-body systems, previous treatments show that three evolutionary paths are available: [A] The orbit of the secondary can decay inward and eventually collide with the primary, where the orbital angular momentum is transferred to the rotation of the primary. [B] The orbit can gain angular momentum from the primary and move outwards toward an unbound state. [C] The system can approach a stable tidal equilibrium configuration, where the orbit and spins of both bodies have the same period and their angular momentum vectors point in the same direction. The Pluto-Charon system provides one such example \citep{tholen}. In order for the equilibrium state to exist, the system must have a minimum amount of total angular momentum; in order for the system to reside in the equilibrium state, its orbital angular momentum must be at least three times larger than the spin angular momentum. The goal of this paper is to derive analogous requirements for the existence of stable tidal equilibrium states for hierarchical three-body systems (star-planet-moon systems). The existence of tidal equilibrium states does not depend directly on the energy dissipation mechanisms that allow systems to attain such states. In the cases of interest, the relevant dissipative evolution is generally driven by tidal effects \citep{hut1981,zahn}. Since tidal evolution occurs over long time scales, often comparable to the age of the universe, many extant systems are not expected to have reached their lowest energy states. Notice also that tidal equilibrium is determined under the assumption of conservation of angular momentum, which must exceed a minimum value for the equilibrium state to exist. Although planetary systems can lose angular momentum via stellar winds and other astronomical processes, systems that start with too little angular momentum generally have no means of gaining more. In the regimes of interest, the host planet often has mass comparable to Jupiter and the moon has mass comparable to Earth. This satellite mass is favored because it is large enough to produce measurable transit timing variations in Hot Jupiter systems and large enough to be potentially habitable in systems with more temperate Jupiters. Here we denote the masses of the three bodies as $M$ for the star, $m$ for the planet, and $\mu$ for the moon. These masses are assumed to obey the ordering \be\mu\ll{m}\ll{M}.\label{ordering}\ee This work assumes that the planetary orbit (about the star) has semimajor axis $a$ and the moon orbit (about the planet) has semimajor axis $b$ (when the moon orbits within a non-Keplerian potential, the scale $b$ refers to the radius of the lunar orbit). The star has spin angular momentum with moment of inertia $I$ and rotation rate $\Omega$. The planet also has spin angular momentum with moment of inertia $J$ and rotation rate $\omega$. Finally, the moon has moment of inertia $K$ and rotation rate $\lambda$. Because of the mass ordering from equation (\ref{ordering}), we can ignore the difference between the reduced mass and the actual mass of an orbiting body. A great deal of previous work has considered the dynamics of these hierarchical systems \citep{szebehely,szebehely78}. However, most of the work regarding system stability does not include the effects of rotation of the bodies (see the review of \citealt{george}). As shown previously for two-body systems, the possible exchange of angular momentum between the orbit(s) and rotation plays an important role in determining tidal equilibrium states. We also note that many previous numerical simulations (e.g., \citealt{payne} and references therein) have shown that the long-term dynamical stability of the system requires the lunar orbit to fall within some fraction $f$ of the Hill radius $(b<fR_H)$, where $R_H\equiv(m/3M)^{1/3}a$ and $f\sim1/2$. These numerical studies generally do not include rotational angular momentum, but provide independent constraints on the stability of star-planet-moon systems. Significantly, the results of this paper indicate that tidal equilibrium states --- when they exist --- correspond to lunar orbits that lie outside the stability regime found numerically. Finally, the tidal evolution of our moon, and others, including tidal dissipation has been studied in detail \citep{goldpeale,touma}). The dissipation time scales often exceed the age of the system, or even the age of the universe, so that observed systems are often not found in their lowest accessible energy states. The goal of this paper is to consider the equilibrium states of the system, including the effects of stellar, planetary, and lunar rotation in the energy and angular momentum budgets. We first consider the problem in the limit where both the orbits can be described using Keplerian solutions (Section \ref{sec:kepler}). For this case, we allow the five angular momentum vectors to have different directions and allow the orbits to have eccentricity. Nevertheless, the tidal equilibrium state corresponds to circular orbits and aligned angular momentum vectors. Moreover, the critical point is a saddle point, rather than a minimum, so that no stable tidal equilibrium state exists for this limiting case. We then generalize the problem to include the stellar influence on the lunar orbit, which is treated by a time-average of the circular restricted three-body problem (Section \ref{sec:beyond}). For this case, the system has two critical points. The first corresponds to a saddle point, whereas the second corresponds to a minimum of the energy so that a stable tidal equilibrium state exists. In Section \ref{sec:apply}, we briefly consider applications of these results for extrasolar planetary systems, as well as moons in our Solar System. The paper concludes, in Section \ref{sec:conclude}, with a summary of our results and a discussion of their implications.

\label{sec:conclude} This paper considers the issue of tidal equilibrium for hierarchical star-planet-moon systems, which are described by 16 variables that specify the energy and angular momentum of the stellar spin, planetary spin, planetary orbit, lunar orbit, and lunar spin. The optimization procedure is subject to the constraint of constant angular momentum. A summary of our results is given below, along with a discussion of their implications. \subsection{Summary of Results} \label{sec:results} The first finding of this paper is that stable tidal equilibrium states do not exist for hierarchical star-planet-moon systems for the case where orbits are described in the Keplerian limit (Section \ref{sec:kepler}). This result stands in contrast to case of binary systems with spin angular momentum, also evaluated in the Keplerian limit, where stable tidal equilibrium states are present as long as the total angular momentum of the system is sufficiently large. In the Keplerian approximation, the system energy has a single critical point which corresponds to synchronous rotation of the star, planet, moon, planetary orbit, and the lunar orbit. In addition, both orbits have zero eccentricity and all five angular momentum vectors are aligned. However, the critical point does not represent a minimum of the energy, but rather a saddle point (Figure \ref{fig:saddle}). Moreover, at the critical point, the lunar orbit lies outside the Hill radius, although this boundary is not defined within the Keplerian approximation. In practice, this finding indicates that the system can evolve to a lower energy state --- while conserving total angular momentum --- by changing the orbital radius of the moon from its critical value. If the moon moves inward toward the planet, orbital angular momentum of the lunar orbit decreases, and the spin and/or the orbital angular momentum of the planet must increase to compensate. If the moon moves outward and increases its orbital angular momentum, the spin and/or orbital angular momentum of the planet decrease accordingly. In the Keplerian limit, the formulation does not include the tidal force from the star acting on the lunar orbit. This interaction term specifies the Hill radius, which provides an effective outer boundary for allowed lunar orbits. One can include this effect in the formulation, provided that the additional terms are time-averaged. We have generalized the calculation so that the lunar orbit is treated using the circular restricted three-body approximation (Section \ref{sec:beyond}). In this case, the energy of the system has two critical points. The first is a saddle point, whereas the second one is a local minimum (Figure \ref{fig:beyond}). As a result, a stable tidal equilibrium state exists for this system. The existence of the tidal equilibrium state requires that the total angular momentum of the system exceeds a well-defined value $L_X$ (given by equation [\ref{lxmin}]), analogous to results obtained previously for binary systems. For $L=L_X$, the combined orbital angular momentum of both orbits is three times the rotational angular momentum of the planet. Even when it exists, the stable tidal equilibrium point lies outside the boundary (roughly half the Hill radius) where lunar orbits are found to be stable in numerical simulations. As a result, the tidal equilibrium point found here does not does not represent a viable long-term state of the system. Although exomoon systems can in principle have stable tidal equilibrium states, most observed systems with exo-jupiters don't have the right angular momentum for a stable tidal equilibrium state to exist even for the star-planet system considered without a moon (as illustrated by Figure \ref{fig:obsplane}). For host stars with measured stellar rotation rates \citep{mcquillan}, the total angular momentum is generally less than the critical value $L<L_X$ and/or the orbital angular momentum is too small relative to the total (see also \citealt{levrard,abtide}). Since these star-planet systems do not reside in tidal equilibrium, existing systems must dissipate energy and evolve. However, the time scales must be longer than typical system ages (several Gyr), and this requirement places constraints on the tidal dissipation parameters. In a similar vein, moons in our Solar System orbit much closer to their host planets than both the tidal equilibrium point and the dynamical stability boundary (Figure \ref{fig:ssmoon}). Although these moons will eventually be exiled as they gain orbital angular momentum from their host planets, they remain in orbit because the Solar System is not old enough for this process to proceed to completion. \subsection{Discussion} \label{sec:discuss} The problem of finding tidal equilibrium states for hierarchical three-body systems is well-defined in principle. However, the application of these results to astronomical systems requires further discussion. First we note that although the full three-body problem has energy and angular momentum integrals \citep{sundman}, the expressions used here for these conserved quantities are necessarily approximate. Because the planetary orbit is nearly Keplerian, the key approximation in this study is the model used to describe the lunar orbit. In general, the angular momentum of the moon (in orbit about the planet) is not constant in time. However, in both the Keplerian limit (Section \ref{sec:kepler}) and the restricted three-body generalization used here (Section \ref{sec:beyond}), the time-averaged lunar orbit does have well-defined angular momentum integral (e.g., see \citealt{goldreich66}). As a result, the constrained optimization procedure of this paper is valid in a time-averaged sense. Since moons are expected to have periods $P\sim1-100d$, whereas evolutionary times scales are $\sim$Gyr, the time-averaged angular momentum of the lunar orbit is indeed well-defined for the time scales over which the system energy is expected to change. Another issue that arises is the long-term dynamical stability of star-planet-moon systems. In the Keplerian limit, the critical point lies outside the Hill radius, and the critical point is a saddle point, so that no chance of stability arises. In the orbit-averaged case (Section \ref{sec:beyond}), the potentially stable critical point lies near the Hill radius that one obtains for the time-averaged problem, and the critical point can be a minimum.\footnote[2]{For completeness, we note that the Hill radius for the time-averaged treatment differs from the Hill radius obtained from the standard circular restricted three-body treatment by a factor of order unity. In the standard case, one obtains $R_h=(m/3M)^{1/3}a$ for the Hill radius, whereas in the time-averaged case, the effective Hill radius becomes $R_h=(2m/3M)^{1/3}a$.} However, numerical simulations show that the stability of lunar orbits requires the semimajor axis to be less than a fraction (typically $f\sim1/2)$ of the standard Hill radius (\citealt{payne} and references therein). In scaled units ($G=M=J=1$ and $b\to{b}m^{-1/3}$), this requirement becomes $b/a<1/(2\cdot3^{1/3})$. In contrast, the value of the ratio $b/a$ at the stable equilibrium point is given by the solution to equation (\ref{xequation}), which implies $b/a\approx(2/3)^{1/3}$. The two locations differ by a factor of $2^{4/3}\approx2.52$. The dynamical stability constraint thus requires the moon to orbit well inside the critical point found here. Such systems can evolve to lower energy states by decreasing the lunar semimajor axis and transferring angular momentum to the planetary orbit and/or the planetary rotational energy. Although star-planet-moon systems often have no stable tidal equilibrium states, the moons in our Solar System exist over Gyr time scales as they evolve \citep{goldsoter}. Possible moons in other systems can have shorter lifetimes \citep{barnes}. Tidal forces act to move the moons outward (inward) if the planetary rotation rate is larger (smaller) than the orbital mean motion. For systems that are dynamically stable, moons must orbit well within the Hill sphere. Since the planetary spin is close to synchronous with the planetary orbit for systems near the critical point, which is near the Hill radius, the lunar mean motion will generally be larger than the planetary spin; in this case, the moon will eventually fall into the planet. For systems residing far from their critical point, however, the planet rotation rate could be super-synchronous, so that the lunar orbit evolves outward and the moon eventually becomes unbound from the planet. We also note that the tidal equilibrium states considered in this paper are related to spin-orbit resonances. Many of the moons in our Solar System are in or near a synchronous spin-orbit resonance where the orbital period of the moon is commensurate with the rotational period of the planet \citep{md99}. These configurations correspond to minimum energy states and planet-moon systems are driven toward such states through the action of dissipative forces. Tidal equilibrium states are also minimum energy configurations. In practice, planet-moon systems in spin-orbit resonance undergo small oscillations about the equilibrium state. In addition, the full description of the equilibrium state includes additional properties of the system, such as the quadrupole moments of the bodies (see Chapter 5 of \citealt{md99}). This paper has only considered the energy states of these hierarchical systems, and not the dynamical evolution toward the equilibrium states. This evolution must take place through dissipative processes such as tidal interaction between the constituent bodies. We note that the time scales for such interactions will generally be different for the star-planet system and the planet-moon system. As a result, one of the two-body subsystems can reach its tidal equilibrium state while the three-body system as a whole remains far from its tidal equilibrium state. In our Solar System, the Sun-Pluto-Charon system provides one such example. The results of this paper change our interpretation of star-planet-moon systems: Lunar orbits in these systems, including our own Solar System, often have no tidal equilibrium states and cannot be absolutely stable (and this result stands in contrast to the case of two-body systems). When stable tidal equilibrium states exist, the required lunar orbits are so distant that dynamical interactions render the systems untenable, so that the moons can be scattered out of their orbits. If moons orbit close enough to their host planets to avoid this fate, they must lie well inside any possible stable equilibrium point, and must evolve through the action of dissipative forces (perhaps on long time scales). As a result, lunar orbits can only persist because the systems are not old enough for them to have dissipated their energy, or for dynamical interactions to scatter the moons out of their planet-centric orbits. \medskip \textbf{Acknowledgments:} We thank Konstantin Batygin, Juliette Becker, Seth Jacobson, Dan Scheeres and Chris Spalding for useful conversations. We also thank an anonymous referee for useful comments.

1607.08170

1607

1607.04218_arXiv.txt

The ultraviolet background (UVB) emitted by quasars and galaxies governs the ionization and thermal state of the intergalactic medium (IGM), regulates the formation of high-redshift galaxies, and is thus a key quantity for modeling cosmic reionization. The vast majority of cosmological hydrodynamical simulations implement the UVB via a set of spatially uniform photoionization and photoheating rates derived from UVB synthesis models. We show that simulations using canonical UVB rates reionize and, perhaps more importantly, spuriously heat the IGM, much earlier $z \sim 15$ than they should. This problem arises because at $z > 6$, where observational constraints are nonexistent, the UVB amplitude is far too high. We introduce a new methodology to remedy this issue, and we generate self-consistent photoionization and photoheating rates to model any chosen reionization history. Following this approach, we run a suite of hydrodynamical simulations of different reionization scenarios and explore the impact of the timing of reionization and its concomitant heat injection on the the thermal state of the IGM. We present a comprehensive study of the pressure smoothing scale of IGM gas, illustrating its dependence on the details of both hydrogen and helium reionization, and argue that it plays a fundamental role in interpreting \lyalpha{} forest statistics and the thermal evolution of the IGM. The premature IGM heating we have uncovered implies that previous work has likely dramatically overestimated the impact of photoionization feedback on galaxy formation, which sets the minimum halo mass able to form stars at high redshifts. We make our new UVB photoionization and photoheating rates publicly available for use in future simulations.

\label{sec:intro} In our current standard model of the universe, hydrogen and helium account for 99\% of the baryonic mass density \citep{Planck:2015}. After the recombination epoch, these elements remain neutral until ultraviolet radiation from star-forming galaxies and active galactic nuclei reionizes them. Therefore, this ultraviolet background (UVB) governs the ionization state of intergalactic gas and plays a key role in its thermal evolution through photoheating. During the reionization of $\HI$ and later $\HeII$, ionization fronts propagate supersonically through the intergalactic medium (IGM), impulsively heating gas to $\sim10^{4}$ K \cite[see, e.g.,][]{Abel:1999,McQuinn:2012,Davies:2016}. As the universe evolves, it is well known that the balance between cooling due to Hubble expansion and inverse-Compton scattering of cosmic microwave background (CMB) photons and heating due to the gravitational collapse and photoionization heating give rise to a well-defined temperature-density relationship in the IGM \citep[][]{Hui:1997,McQuinn:2012}: \begin{equation} T=T_{0}\times \Delta^{\gamma-1} \label{eq:T0gamma} \end{equation} where $\Delta=\rho/\bar{\rho}$ is the overdensity with respect to the mean and $T_{0}$ is the temperature at the mean density. Immediately after the reionization of $\HI$ ($z \lesssim 6$) or $\HeII$ ($z \lesssim 3$), $T_{0}$ is likely to be around $\sim2\times 10^{4}$ K and $\gamma\sim1$ \citep{Bolton:2009,McQuinn:2009}; at lower redshifts, $T_{0}$ decreases as the universe expands, while $\gamma$ is expected to increase and asymptotically approach a value of $1.62$ \citep{Hui:1997}. Another important physical ingredient to describe the thermal state of the IGM is the gas pressure support. At small scales and high densities, baryons experience pressure forces that prevent them from tracing the collisionless dark matter. This pressure results in an effective 3D smoothing of the baryon distribution relative to the dark matter, at a characteristic scale. known as the Jeans pressure smoothing scale, $\lj$. In an expanding universe with an evolving thermal state, this scale at a given epoch is expected to depend on the entire thermal history, because fluctuations at earlier times expand or fail to collapse depending on the IGM temperature at that epoch \citep{Gnedin:1998,Kulkarni:2015}. Recently, \citet{Rorai:2013} and \citet{Rorai:2015} have shown that an independent measurement of the pressure smoothing scale can be obtained using the coherence of \lyalpha{} forest absorption in close quasar pairs \citep{Hennawi:2006,Hennawi:2010}. \lyalpha{} forest observations between $2<z<6$ probe the moderate overdensities characteristic of the IGM and therefore are a crucial tool to understand the properties of the UVB. In the last decade, the precision of these measurements has continued to grow both in terms of their numbers (BOSS\footnote{Baryon Oscillation Spectroscopic Survey (BOSS): https://www.sdss3.org/surveys/boss.php} survey) and in quality \citep[high signal-to-noise ratio spectrum from, e.g.][]{OMeara:2015}. However, while it seems that we keep learning more and more about the ionization history of the universe, for both $\HI$ and $\HeII$ reionizations \citep[e.g.][]{Becker:2013,Syphers:2014,Worseck:2014,Becker:2015} the thermal history of the universe is still far from certain. The statistical properties of the \lyalpha{} forest are sensitive to the thermal state of the gas, trough both thermal broadening of lines and pressure support. When constraints on the thermal history are reviewed, they yield very puzzling results. Measurements of $T_{0}$ from different groups utilizing different methodology are in poor agreement \citep{Schaye:2000,Bolton:2008,Lidz:2010,Becker:2011, Rudie:2012,Garzilli:2012,Boera:2014,Bolton:2014}. A similar problem appears when measurements of the slope of the temperature–density relation, $\gamma$, are compared. At $z\simeq3$ some authors have even found that $\gamma$ is either close to isothermal ($\gamma=1$) or even inverted \citep[$\gamma<1$;][but see \citealt{Lee:2015}]{Bolton:2008,Viel:2009}. Most studies of the thermal state of the IGM ignore uncertainties resulting from the unknown pressure smoothing scale \citep[but see][]{Becker:2011,Puchwein:2015}, which produces a 3D smoothing that is difficult to disentangle from the the similar but 1D smoothing resulting from thermal broadening \citep{Peeples:2010a,Peeples:2010b,Rorai:2013}. Therefore, ignoring this effect has probably contributed to the confusing and sometimes contradictory published constraints on $T_{0}$ and $\gamma$ \citep{Puchwein:2015}. With the help of accurate models of the IGM, the statistics of the \lyalpha{} forest can be used to constrain its thermal parameters and ultimately cosmic reionization. Ideally one will run coupled radiative transfer hydrodynamical simulations that include extra physics governing the sources of ionizing photons (stars, quasars, etc.). Despite significant progress on this front \citep{Wise:2014,So:2014,Gnedin:2014,Pawlik:2015,Norman:2015,Ocvirk:2015} these simulations are still too costly for sensible exploration of the parameter space. For this reason, the dominant approach, implemented in the vast majority of hydrodynamical codes, is to assume that all gas elements are optically thin to ionizing photons, such that their ionization state can be fully described by a uniform and isotropic UV+X-ray background radiation field. Thus, the radiation field is encapsulated by a set of photoionization and photoheating rates that evolve with redshift for each relevant ion. The minimal set of ions are $\HI$, $\HeI$ and $\HeII$ in order to track the most relevant ionization events, as well as the thermal heating associated with them. Of course, although this optically thin approximation is a valid assumption once the mean free path of ionization photons, $\lambda_{\rm mfp,\nu}$, is large enough, it is certainly not true during cosmic reionization events. As such, this optically thin approach is not meant to provide an accurate description of reionization itself, but it should at least provide a reasonable description of the heat injection associated with reionization. This is important since galaxies forming during the reionization epoch are sensitive to the thermal state of the gas, and even well after reionization gas elements can retain thermal memory of reionization heating \citep{Gnedin:1998, Kulkarni:2015}. It is important to remark here that these UVB models have relevant consequences for galaxy formation and evolution models and hydrodynamical simulations. Several groups have already shown how important the UVB model is to determine the star formation of the first galaxies and their evolution by not only setting the minimum halo mass able to form stars \citep[i.e., halos massive enough to overcome gas pressure forces;][]{Rees:1986,Sobacchi:2013} but also regulating the gas accretion from the IGM into the more massive halos \citep{Quinn:1996,Simpson:2013,BenitezLlambay:2015,Wheeler:2015a}. The standard approach is to adopt photoionization and photoheating rates from semianalytical synthesis models of the UVB \citep{Haardt:1996,Haardt:2001,FaucherGiguere:2009,Haardt:2012}. However, these UVB synthesis models surely break down during reionization events, and the validity of using them in optically thin simulations (during reionization) is questionable. Moreover, as we will show, these models are fundamentally inconsistent during reionization, leading to different reionization histories in the simulations than the ones given by the authors. Specifically, they reionize the universe too early, and as a result they produce spurious heating of the IGM at early times (see Section~\ref{sec:typmodels} and Figure~\ref{fig:Qhistgas0}). In this paper, we improve on the limitations of current UVB models to provide reliable ionization and thermal histories during reionization by developing a new method to model ionization and heating during reionization in hydrodynamical simulations. In the context of this method, we demonstrate how to run simulations with self-consistent ionization and thermal histories that agree with constraints from the CMB and IGM measurements. Moreover, we make these new tables publicly available in the default format used by most cosmological codes. \begin{figure*} \begin{center} \includegraphics[angle=0,width=\textwidth]{QtauT0_hist0} \caption{Ionization and thermal history of the universe obtained in hydrodynamical simulations using standard tables from \citet{Haardt:2012}. The upper panel shows the evolution of the $\HII$, $\HeIII$ volume-averaged ionized fractions in the simulation. Dashed lines stand for the $\mean{x_{\HII}}$ and dot-dashed lines for the $\mean{x_{\HeIII}}$) The full and the dot-dashed black lines stand for the $Q_{\HII}$ (solid lines), $Q_{\HeIII}$ (dot-dashed lines) volume filling factors calculated by \citet{Haardt:2012} for their model. The middle panel shows the integrated electron scattering optical depth, $\taue$. The gray band stands for \citet{Planck:2015} constraints on $\taue$ coming from the CMB. The lower panel shows the evolution of the temperature at mean density. Notice that reionization finishes much earlier in the simulation. All parameters presented in this figure are converged within $<5\%$ accuracy (see Section~\ref{ssec:convergence}). \label{fig:Qhistgas0}} \end{center} \end{figure*} The outline of the paper is as follows. In Section~\ref{sec:typmodels} we discuss in detail current standard methods that include the effect of the UVB in optically thin hydrodynamical simulations. We show that these models have problems reproducing the desired ionization and thermal histories. In Section~\ref{sec:newmodel} we present a new method to improve the current models of the UVB during reionization events. The different reionization models considered in this work, based on current observational constraints, are motivated in Section~\ref{sec:reionmodels}. We describe the basic details of the hydrodynamical cosmological code that we have used in this work, the analysis pipeline, and the properties of the simulations in Section~\ref{sec:code}. The ionization and thermal histories of the simulations using the new UVB models are shown and examined in Section~\ref{sec:results}. We explore the possibility of reproducing observational constraints on the $\HI$ and $\HeII$ transmission in Section~\ref{sec:meanflux}. In Section~\ref{sec:discuss} we discuss the limitations of our new approach, provide a comparison to previous work, and discuss previous work using incorrect UVB models in galaxy formation simulations that likely overestimate the impact of photoionization feedback. We conclude in Section~\ref{sec:conc}. In Appendix~\ref{app:volave} we provide details on how the new photoionization and photoheating rates are derived in our method. In Appendix~\ref{app:HMold} we present the ionization and thermal histories of several widely used UVB models. The effects of cosmology on the new models are discussed in Appendix~\ref{app:cosmo}. Finally, in Appendix~\ref{app:tables} we present the photoionization and photoheating rates of the new models.

\label{sec:conc} In this paper we have presented results from optically thin cosmological hydrodynamical simulations using the Nyx code \citep{Almgren:2013,Lukic:2015}. As commonly done in multiple IGM and galaxy formation studies, the UV background is modeled as a uniform and isotropic field that evolves with redshift. Operationally, the UVB determines the photoioinization and photoheating rates of $\HI$, $\HeI$ and $\HeII$, which are inputs to the code. We have demonstrated that when canonical models of the UVB, like that of HM12, are used in hydrodynamical simulations, the ionization of the IGM and, more importantly, the concomitant heating occur far too early, inconsistent with the reionization histories calculated by the respective authors, and in violation of current observational constraints on reionization. We argue that this results from the fact that these models dramatically overestimated the mean free path of ionizing photons at root at $z > 5$, resulting from the blind extrapolation of a model fit to lower redshift ($z < 5$) measurements. As a result, the amplitudes of the photoionization and photoionization heating rates are far too high at $z > 6$. This premature heating spuriously heats the IGM to $\sim 10^4$ K by $z \sim 13$, and because the IGM gas pressure smoothing scale depends on the full thermal history, it produces an erroneously large pressure smoothing scale at nearly all redshifts. We argue that a correct and consistent model of the reionization and thermal history is crucial for obtaining the correct pressure scale in simulations, which is necessary for interpreting \lyalpha{} forest statistics at $z<6$ -- not doing so can bias estimates of the thermal state of the IGM. We also discussed the implications of this spurious early heating on galaxy formation simulations. Motivated by these issues, we have developed a new method to generate UVB models for hydrodynamical simulations that allow one to self-consistently simulate different reionization models. We implement this by volume-averaging the photoionization and energy equations. In the model, each reionization event is defined by the ionization history with redshift and the total heat input of the reionization event. In this sense our new models provide a very promising tool to explore the parameter space of possible ionization and thermal histories. In this work, we investigated models in which we changed the redshift at which $\HI$ reionization ends and the amount of heat input associated with both $\HI$ and $\HeII$ reionization. We studied the effect of these changes on the thermal history of the IGM, in particular the temperature at mean density, $T_{0}$, the slope of the temperature-density relation, $\gamma$, the temperature at the optimal density probed by curvature measurements, $T(\Deltaopt)$, and the gas pressure smoothing scale, $\lambda_{P}$. We have shown how important the degeneracies between these parameters can be in order to derive the thermal parameters of the IGM using the curvature \lyalpha{} statistic. These rates have also been corrected to improve the agreement with measurements of the average $\HI$ and $\HeII$ transmission after reionization. The UVB plays a fundamental role in determining the star formation of the first galaxies and their evolution by not only setting the minimum halo mass able to form stars but also regulating the gas accretion from the IGM into more massive halos. Previous studies utilizing UVB models that suffer from the spurious early heating described in this paper have thus overestimated the effect of this photoheating feedback and the resulting suppression of star-formation at high redshift. We therefore argue that galaxy formation simulations should be revisited using our new UVB models. We make our new UVB models publicly available so that the community can better explore the consequences and effects of different ionization and thermal histories in all types of hydrodynamical cosmological simulations. Tables with the photoionization and photoheating rates of the new models can be found in Appendix~\ref{app:tables}, which are in the the standard ``TREECOOL'' file format, ready and easy to use with most cosmological codes, including \textsc{gadget}, \textsc{arepo}, and \textsc{gizmo}. We encourage anyone interested in implementing some other specific UVB model to contact the authors. We will also be happy to provide help incorporating these models into other codes by request. Like all optically thin simulations, our approach misses UVB fluctuations that could produce scatter in reionization times and temperatures between different regions of the universe \citep[e.g.][]{Abel:1999,Meiksin:2004,Pontzen:2014,GontchoaGontcho:2014,Davies:2014,Malloy:2015,DAloisio:2015,Davies:2016b}. Of course, an immediate solution to this problem in optically thin simulations will be to add some extra dependence on a specific property of each resolution element in the simulation (e.g. density, distance from a halo that could host a galaxy/quasar, etc). However, the computational challenge is to try to do this without making the simulation prohibitively expensive. In this sense, an interesting solution that it is worth exploring will be to assign a specific reionization redshift to each resolution element of the simulation using, for example, an excursion set formalism \citep[e.g.][]{Furlanetto:2004}. In this picture, a resolution element will not see the UVB background until its reionization redshift. We plan to pursue this idea in the near future. Another very valuable piece of information that would improve current optically thin hydrodynamical simulations would be if future radiative transfer simulations \citep[e.g.][]{So:2014,Gnedin:2014,Pawlik:2015,Norman:2015,Ocvirk:2015} were to make the probability distribution function of their photoionization rates publicly available (and perhaps also its dependence on density), and not just the evolution of the mean and/or median values. Finally, our new parameterization for the heating and ionization produced by the UVB allows us to explore a broader range of reionization models, as well as any other physical scenarios that could alter the thermal history of the IGM. This will allow us to better test the effect of such models in simulations of galaxy formation and the IGM. These models could include Population III stars \citep{Manrique:2015}, X-ray pre-heating coming from from starburst galaxies, supernova remnants, or miniquasars \citep{Oh:2001,Glover:2003,Madau:2004,Furlanetto:2006}, dark matter annihilation or decay \citep[][]{Liu:2016} or cosmic rays \citep{Samui:2005}, from the intergalactic absorption of blazar TeV photons \citep{Chang:2012,Puchwein:2012}, or from broadband intergalactic dust absorption \citep{Inoue:2008}. We expect more detailed studies on these physically motivated models in the future.

1607.04218

1607

1607.03438_arXiv.txt

We revisited the spectroscopic characteristics of narrow-line Seyfert 1 galaxies (NLS1s) by analysing a homogeneous sample of 296 NLS1s at redshift between 0.028 and 0.345, extracted from the Sloan Digital Sky Survey (SDSS-DR7) public archive. We confirm that NLS1s are mostly characterized by Balmer lines with Lorentzian profiles, lower black hole masses and higher Eddington ratios than classic broad-line Seyfert 1 (BLS1s), but they also appear to be active galactic nuclei (AGNs) contiguous with BLS1s and sharing with them common properties. Strong Fe\,{\sc ii} emission does not seem to be a distinctive property of NLS1s, as low values of Fe\,{\sc ii}/H$\beta$ are equally observed in these AGNs. Our data indicate that Fe\,{\sc ii} and Ca\,{\sc ii} kinematics are consistent with the one of H$\beta$. On the contrary, O\,{\sc i}\,$\lambda$8446 seems to be systematically narrower and it is likely emitted by gas of the broad-line region more distant from the ionizing source and showing different physical properties. Finally, almost all NLS1s of our sample show radial motions of the narrow-line region highly-ionised gas. The mechanism responsible for this effect is not yet clear, but there are hints that very fast outflows require high continuum luminosities ($>10^{44}$ erg s$^{-1}$) or high Eddington ratios ($\log(L_{\rm bol}/L_{\rm Edd})>-0.1$).

Known for thirty years \citep{1985ApJ...297..166O}, narrow-line Seyfert 1 galaxies (NLS1s) were conventionally classified as those Seyfert 1 galaxies having full width at half maximum (FWHM) of H$\beta$ lower than 2000 km s$^{-1}$ \citep{1989ApJ...342..224G} and showing the flux ratio between [O\,{\sc iii}]$\lambda$5007 and the total (broad plus narrow) H$\beta$ smaller than 3 \citep{1985ApJ...297..166O}. Most of them are also characterized by emission of strong Fe\,{\sc ii} multiplets, [Fe\,{\sc vii}]$\lambda$6087 and [Fe\,{\sc x}]$\lambda$6374, and in some cases also by [Fe\,{\sc xi}]$\lambda$7892 and [Fe\,{\sc xiv}]$\lambda$5303 \citep{2008MNRAS.385...53M, 2011nlsg.confE...2P}. In addition, NLS1s show a steep soft X-ray slope \citep{1996A&A...305...53B}, rapid and large soft X-ray variability \citep{2004MNRAS.347..269G}, a weak big blue bump in the optical/UV range \citep[e.g.][]{1994A&A...283L...9M, 1995MNRAS.276.1281P}, bright IR emission \citep{1996ApJS..106..341M}, and nuclear super-solar metallicity \citep[and references therein]{2000MNRAS.314L..17M, 2001A&A...374..914K}. Some NLS1s show also radio and gamma-ray emission, signs of the presence of a relativistic jet \citep{2009ApJ...707L.142A, 2015A&A...575A..13F}. NLS1s are generally believed to be powered by less massive black holes (BHs) than classic broad-line Seyfert 1 galaxies (BLS1s) \citep{1996A&A...305...53B, 2000NewAR..44..491P} and accreting up to the Eddington limit \citep{1992ApJS...80..109B}. Therefore, it is likely that NLS1s are a young phase of BLS1s \citep{2000MNRAS.314L..17M}. Notwithstanding all these established properties, some questions still remain unsolved. Models by \citet{2012MNRAS.426.3086G} suggest that NLS1s are expected to have broad-line region (BLR) lines with Lorentzian profiles caused by a macroscopic turbulent motion of gaseous clouds, which, especially at large radii, contributes significantly to the line wings. At the same time, \citet{2011Natur.470..366K, 2013A&A...549A.100K} showed that the profiles of permitted lines in active galactic nuclei (AGNs) consist mainly of rotationally broadened Lorentzian functions. In the past, other authors favored a fitting with a single Lorentzian function \citep[see e.g.][]{1996ApJS..106..341M, 2001A&A...372..730V, 2002ApJ...566L..71S}. However, there is no consensus on this topic, also because it is still debated whether a Lorentzian function is the most appropriate representation of the line profile, since a double Gaussian, or even a multi-Gaussian approach seemed to provide a more statistically robust fitting \citep[see e.g.][]{2005ApJ...623..700D, 2008MNRAS.385...53M}. Strong multiplets of Fe\,{\sc ii} are often detected in NLS1s spectra. The excitation mechanism of these spectral lines is still uncertain. Photoionization models cannot easily account for strong Fe\,{\sc ii} emission \citep{2000NewAR..44..531C}, since the region emitting these multiplets is a weakly ionized part of the BLR. Assuming a high over-abundance of iron also does not reproduce the intensity of the multiplets \citep{2000NewAR..44..531C}. Shocks induced by winds and/or outflows have been invoked to heat the gas and cause the collisional excitation of the upper energy levels of the Fe\,{\sc ii} transitions \citep{2000NewAR..44..531C, 2006A&A...451..851V, 2010ApJ...721.1835S}. Fe\,{\sc ii} emission could depend on the physical conditions of the low-ionization clouds of the BLR which emits also O\,{\sc i}\,$\lambda$8446, and Ca\,{\sc ii}\,$\lambda$8498,8542,8662 (hereafter CaT). Indeed, the ionization potentials of O\,{\sc i} and Ca\,{\sc ii} (13.6 and 11.9 eV, respectively) are close to that of Fe\,{\sc ii} (16.2 eV). These lines are not yet extensively studied likely because of their proximity to the NIR domain, which requires spectroscopic infrared observations when the redshift is larger than 0.1. CaT is a transition onto a metastable triplet produced by collisional excitation and likely emitted by a high-density, optically-thick neutral gas \citep{1988ApJ...330..751P, 2007ApJ...663..781M, 2013AstRv...8d...4M, 2015ApJS..217....3M}. O\,{\sc i}\,$\lambda$8446 can be emitted in less extreme conditions \citep{2013AstRv...8d...4M, 2015ApJS..217....3M}, but its origin is not yet completely understood. The O\,{\sc i}\,$\lambda$8446 line is generally believed to be produced by Ly-$\beta$ fluorescence, that excites electrons from level 2p$^4$ 3P to the higher level 3d 3D$^0$. The other possible physical processes are collisional excitation and recombination. Their contribution to O\,{\sc i}\,$\lambda$8446 can be evaluated on the basis of the presence and intensity of O\,{\sc i}\,$\lambda$7774. The ratio 7774/8446 is expected to be $\sim$0.3 in case of collision \citep{1980ApJ...238...10G}, while it should be $\sim$1.1 in case of recombination \citep{2008ApJS..174..282L}. A comparison among the kinematics of Fe\,{\sc ii}, O\,{\sc i} and CaT in UV, optical and NIR has been recently used to discuss the site of the Fe\,{\sc ii} emitting region, which seems to be the same of O\,{\sc i} and CaT \citep{2016arXiv160205159M} and located in an outer part of the BLR with respect to hydrogen \citep{2015ApJS..221...35K}. On the contrary, reverberation-mapping analyses of a small number of AGNs suggest that both Fe\,{\sc ii} and hydrogen lines are produced by the same ionized gas \citep{2015ApJ...804..138H}. Seyfert 1 galaxies, and NLS1 in particular, are known to show often blue asymmetric profiles in their [O\,{\sc iii}]$\lambda\lambda4959,5007$ emission lines \citep{1981ApJ...247..403H,1985ApJ...294..106V,1985MNRAS.213....1W,1991ApJS...75..383V}. These profiles consist mainly of two components: 1) a narrow \textit{core} component (FWHM$\sim$200--500 km\,s$^{-1}$), believed to be emitted by gas located far from the active nucleus and whose kinematics is dominated by the gravitational potential of the host galaxy, as shown by \citet{1996ApJ...465...96N}; 2) a broad component (FWHM$\sim$500--1000 km\,s$^{-1}$), called \textit{wing}, believed to be emitted by gas located closer to the active nucleus and whose width is dominated by turbulent outflows in a direction perpendicular to the accretion disk \citep{2001A&A...372..730V, 2005MNRAS.364..187B, 2008ApJ...680..926K, 2011ApJ...737...71Z}. Furthermore, in some AGNs the high ionization emission lines show a systematic difference in velocity when compared to the low ionization lines. First detected by \citet{1976ApJ...208...37P} in I\,Zw\,1, this effect was studied more recently by several authors \citep[see e.g.][]{2002ApJ...576L...9Z,2003MNRAS.345.1133M,2005ApJ...618..601A,2005AJ....130..381B,2005MNRAS.364..187B,2008ApJ...680..926K,2011ApJ...737...71Z,2015arXiv151107138M,2015arXiv151202642Z}. AGNs with [O\,{\sc iii}] showing a velocity shift of $\Delta v<-250$ km s$^{-1}$ with respect to H$\beta$ were called \textit{blue outliers} by \citet{2002ApJ...576L...9Z}. More recently, \citet{2008ApJ...680..926K} applied a less-restrictive upper limit of $-150$ km s$^{-1}$. This blueshift could be due to an outflow associated with a disk wind \citep{2002ApJ...576L...9Z}, or with a hot decelerating wind \citep{2008ApJ...680..926K} combined with a source of opacity. \citet{1982ApJ...263...79G} found that also the high-ionization lines of BLR, and in particular C\,{\sc iv} line, are blueshifted by about 600 km\,s$^{-1}$ with respect to the rest-frame. A recent work by \citet{2013ApJ...769...30G} points out how much these questions about radial motions of ionized gas in AGNs are still open. On the basis of velocity--resolved reverberation mapping, these authors suggested that blueshifting in BLR could be due to Rayleigh scattering associated to inflowing gas towards the super-massive BH (SMBH). We addressed these open questions by selecting a sample of NLS1s and we focused our work on the following topics: 1) the analysis of the shape of the BLR H$\beta$ emission line to quantify how much a Lorentzian profile is typical in these AGNs and to investigate if the width of H$\beta$ is connected to other physical parameters of the active nucleus like the Eddington ratio; 2) the analysis of the BLR emission lines from Fe\,{\sc ii}, O\,{\sc i} and Ca\,{\sc ii} to clarify if a strong Fe\,{\sc ii} emission is a distinctive property of NLS1s and to explore the physical properties of the low ionization gas; 3) the analysis of the radial motions of ionized gas observed in the narrow-line region (NLR) to understand which are the main drivers and whether the blue wings are related to the shift of the core component. In Sections 2 and 3 we describe the selection of the sample of NLS1--candidate galaxies with its main properties, and the details of the spectroscopic analysis. The results of the analysis are shown in Section 4 and discussed in Section 5. In Section 6, we summarize the results of this work. We adopt a cosmology with $H_0=70$ km s$^{-1}$ Mpc$^{-1}$, $\Omega_M=0.3$ and $\Omega_{\Lambda}=0.7$ to compare our results with the most recent papers.

\subsection{H$\beta_{\rm b}$ profile} The question of the shape of the permitted emission lines in NLS1s is still open. NLS1s seem to have symmetric emission lines with Lorentzian rather than Gaussian shapes. Indeed, many authors applied a Lorentzian function to fit the profile of H$\beta$ \citep{1996ApJS..106..341M, 1999ApJS..125..317L, 1999A&A...341..662G, 2001A&A...372..730V, 2006ApJS..166..128Z, 2012AJ....143...83X}. \citet{1999ApJS..125..317L} affirmed that a Lorentzian profile in NLS1s provides a substantially better description of the H$\beta$ shape than a single Gaussian. Lorentzian profiles give, as a result, narrow emission line ratios in good agreement with those observed in the NLR of Seyfert 2 galaxies as found by \citet{2001A&A...372..730V} who suggested also that the broad Balmer lines could be emitted by a large disk. \citet{2012AJ....143...83X} affirmed that a single Lorentzian profile is essentially equivalent to a double-Gaussian profile when fitting NLS1s, while it is not appropriate to fit BLS1s. On the contrary, \citet{2005ApJ...623..700D} found that Lorentzian profiles yield less satisfying results than employing two Gaussian components having different FWHMs. \citet{2007ApJ...654..754N} used two broad Gaussians to fit the H$\beta$ profile, while \citet{2008MNRAS.385...53M} and \citet{2011ApJ...737...71Z} applied a multi-Gaussian fitting. The double Gaussian can be justifed by assuming that the BLR is made of two kinematically different regions placed at different distances from the central source: an intermediate-line region and a very broad-line region \citep[see e.g.][and references therein]{2010ApJS..189...15K}. A multi-Gaussian approach could suggest that the BLR is made of a continuum of different regions. Our analysis of the H$\beta$ broad emission line in a sample of 296 NLS1s and 64 BLS1s suggests that the Lorentzian profiles are predominant in NLS1s. More in general, a Lorentzian profile seems to be more suited for objects having ${\rm FWHM}<4000$ km~s$^{-1}$ \citep{2002ApJ...566L..71S}. Our BLS1s appear to be transition objects between NLS1s and the BLS1s family sharing common properties with NLS1s, in agreement with \citet{2000ApJ...536L...5S} and \citet{2000NewAR..44..503L} who declared that NLS1s show a clear continuity with BLS1s and that they do not really represent a disjoint class of AGNs. Furthermore, our upper limit to define a NLS1 (FWHM(H$\beta_{\rm b}$) $<$ 2200 km s$^{-1}$) depends on the spectral resolution of our data rather than on a different distribution of FWHM for the two samples. Therefore, it seems that the main difference between the BLS1 and the NLS1 families is not the widths of the broad line profiles, but their shapes. Indeed, in our sample there is a continuity in the distribution of profiles going from the Lorentzian symmetric shapes of almost all NLS1s (with the exception of 5 cases) to the double-Gaussian and often asymmetric profiles of about a third of our BLS1s. This effect can be explained if we assume that the BLR motion is due to a turbulent component kinematically represented by a Lorentzian profile combined with a rotational component represented by a Gaussian profile \citep{2013A&A...549A.100K}. NLS1s would be slow-rotating AGNs, in which narrower Lorentzian profiles are the effect of the dominant role of the turbulent motions \citep{2012MNRAS.426.3086G}. This, in turn, can be due to the presence of less massive BHs in NLS1s than in BLS1s, even if an inclination effect of a flat distribution of BLR clouds cannot be completely ruled out \citep{2008MNRAS.386L..15D} and it could explain the spread in the FWHM distribution, as suggested by \citet{2014Natur.513..210S}. However, other authors \citep[see e.g.][]{2005MNRAS.356..789B, 2005ApJ...627..721G, 2007ApJ...667L..33K, 2011ApJ...739...28X, 2015ApJ...801...38W} have shown that lower BH masses in NLS1s are associated to lower stellar velocity dispersion. This is in agreement with the results of \citet{2015A&A...575A..13F} and \citet{2015A&A...578A..28B}, who measured the BH masses with the second-order momentum of the line, which is independent of the inclination. Therefore, we do not believe that, in general, NLS1s show narrower permitted lines because of an inclination effect of a disk-like BLR. Lower BH masses imply higher normalized accretion rates in NLS1s with respect to BLS1s, as confirmed also by our data. Indeed, the objects fitted by a Lorentzian function show also higher values of Eddington ratio than the objects fitted by a double-Gaussian profile. Since they are mostly NLS1s, this result reinforces the idea that NLS1s are characterized by both Lorentzian shape profiles and high Eddington ratios. \citet{2007ApJ...654..754N} suggested that the Eddington ratio could be used to define NLS1s in place of the classic criterion of FWHM(H$\beta)<2000$\,km\,s$^{-1}$. In particular, a ratio $\frac{L_{\rm bol}}{L_{\rm Edd}}\geq 0.25$ should imply NLS1s. Before comparing physical quantites from different works, it is necessary to take into account that the values of the Eddington ratios depend significantly on the methods used to measure the BH mass and to calculate the bolometric luminosity. Since \citet{2007ApJ...654..754N} followed a different approach, we re-calculated both our BH masses and Eddington ratios by means of their formulas, finding, on average, BH masses 0.2 dex higher and Eddington ratios $0.37$ dex lower. This last value moves the proposed criterion to $\frac{L_{\rm bol}}{L_{\rm Edd}}\geq 0.6$. The median FWHM(H$\beta$) is about 2000 km s$^{-1}$ around $\frac{L_{\rm bol}}{L_{\rm Edd}} = 0.6$ and 261 out of 296 NLS1s are above this limit, while 54 out of 64 BLS1s are below. This confirms that, also in our case, the Eddington ratio can be effectively used to classify NLS1s. \subsection{Fe\,{\sc ii} properties} NLS1s are known to be strong Fe\,{\sc ii} emitters. Indeed, the highest values of R4570 are observed in objects having FWHM(H$\beta)< 4000$ km~s$^{-1}$ \citep{2014Natur.513..210S}. However, it is also true that NLS1s show a wide range of R4570 values, therefore this cannot be a distinctive property of these AGNs. \citet{2006ApJS..166..128Z}, \citet{2007ApJ...654..754N} and \citet{2014Natur.513..210S} calculated the ratio between Fe4570 and H$\beta$ broad and obtained ranges extended up to more than 2, as it can be seen in Fig.~\ref{R4570_FW_Hb}, where we plotted R4570 versus FWHM(H$\beta_{\rm b}$) finding that our data are confined to a smaller range of R4570 values compared to those from \citet{2006ApJS..166..128Z} (green plus symbols) and \citet{2014Natur.513..210S} (black contours). \begin{figure} \includegraphics[width=\columnwidth]{fig29.eps} % \caption{The values of R4570 are plotted against the FWHM of the broad component of H$\beta$. Solid circles are NLS1s, while open circles are BLS1s. Green plus symbols are the data from \citet{2006ApJS..166..128Z}, while the contours are adapted from fig.~1 of \citet{2014Natur.513..210S}. } \label{R4570_FW_Hb} \end{figure} The data from \citet{2007ApJ...654..754N} were not available. It is difficult to give a convincing answer to this difference. First of all, we must remind that we used the total flux of H$\beta$, however the negligible contribution of the narrow component does not affect the following analysis. Our sample was matched with the \citet{2006ApJS..166..128Z} one, finding 120 objects in common. If we consider only this sample, the resulting distribution is consistent with ours: the median value of R4570 measured by Zhou et al. is 0.54, perfectly in agreement with the median value of our sample of NLS1s. Both \citet{2006ApJS..166..128Z} and \citet{2014Natur.513..210S} chose an upper limit in redshift higher than ours, $z < 0.8$ and $z < 0.9$ respectively, therefore we checked whether the luminosity could cause this difference. \citet{2004ApJ...614..558N} suggested that Fe\,{\sc ii}/H$\beta$ depends on the luminosity and, in particular, that the most extreme values of this ratio require very high luminosity. Our values show a range similar to those of \citet{2006ApJS..166..128Z} and \citet{2014Natur.513..210S}, $\lambda L_{\lambda}$(5100) $\sim 10^{42.5-45}$ and $\sim 10^{44-45.5}$ respectively. By plotting R4570 versus $\lambda L_{\lambda}$(5100) we observed that our measurements are included within those of Zhou et al., and, mostly, that their higher values of R4570 belong to the upper end of their luminosity range, but they are not the highest ones (Fig.~\ref{R4570_L5100}). Therefore, a bias in luminosity can be excluded. We estimated both EW(H$\beta$) and EW(Fe4570) from the measurements of \citet{2006ApJS..166..128Z} and \citet{2014Natur.513..210S}. The K--S test shows that the two distributions are significantly different from ours (with p-values $<< 10^{-7}$). Indeed, there is a higher fraction of objects having low EW(H$\beta$) and high EW(Fe4570) in the sample of Zhou et al. and Shen \& Ho, which could justify the excess of high R4570 ratios. Rejecting the hypothesis that our limited range of R4570 values is caused by our lower limit of S/N ratio, which should affect the sample in the opposite direction, the last hypothesis is the less constrained selection criteria applied by \citet{2006ApJS..166..128Z} and \citet{2014Natur.513..210S}, which allowed them to extract and then analyse samples 10--100 times larger than ours. \begin{figure} \includegraphics[width=\columnwidth]{fig30.eps} % \caption{The values of R4570 are plotted against the $\lambda L_{\lambda}$(5100). Symbols are the same as in Fig.~\ref{R4570_FW_Hb}, except for black dots, that are data from \citet{2011ApJS..194...45S}.} \label{R4570_L5100} \end{figure} \citet{2004ApJ...614..558N} found that R4570 correlates with $L/L_{\rm Edd}$ ($r_{\rm s}=0.48$), and suggested that the increase of R4570 with the Eddington ratio is likely caused by the decrease of EW(H$\beta$) with $L/L_{\rm Edd}$. Our data confirm this correlation: our NLS1s show larger ranges of both R4570 and Eddington ratio, than our BLS1s which are confined to lower values. We agree with their hypothesis, furthermore our data suggest that R4570 correlates simultaneously with the equivalent widths of Fe\,{\sc ii} ($r_s = 0.64$) and H$\beta$ ($r_s=-0.60$) at a high level of significance. This is in agreement with similar previous findings by \citet{1992ApJS...80..109B} and \citet{1996A&A...309...81W}. The Eddington ratio is also considered the main driver to explain the R4570--EW([O\,{\sc iii}]) anti-correlation and, more generally, all the correlations forming the so-called EV1 \citep[see e.g.,][]{1994ApJ...435..611L,2000NewAR..44..503L,2000ApJ...536L...5S,2001ApJ...558..553M,2002ApJ...565...78B,2004MNRAS.350L..31B,2004AJ....127.1799G,2004ApJ...614..558N,2009ApJ...703L...1D}. On the other hand, other physical properties have been proposed as underlying drivers, like the black hole mass and inclination angle \citep[see e.g.,][]{1994ApJ...435..611L,1996A&A...309...81W}. The orientation hypothesis was already rejected by some authors \citep[see e.g.,][and references therein]{2000ApJ...542..631K,2004AJ....127.1799G}. The idea that the Eddington ratio is the most important factor was later confirmed and re-inforced by \citet{2014Natur.513..210S}, who analysed about 20 000 broad line AGNs extracted from the SDSS. EV1 was successfully tested with our data: R4570 and EW([O\,{\sc iii}]) are well and significantly anti-correlated. The analysis was extended and the [O\,{\sc iii}] core and wing components were separated. Clear anti-correlations were found, but in case of the wing component the scatter is larger. Similarly, \citet{2014Natur.513..210S} decomposed the [O\,{\sc iii}]$\lambda$5007 profile by fitting each line with core plus wing components, but they found that while the core component anti-correlates strongly with R4570, the same does not occur for the wing component. They argued that the excitation of the core component is dominated by photoionization, as expected, while the outflowing gas of the wing component is likely excitated by additional mechanisms, as shocks. Finally, it is worth to discuss the existence of the anti-correlation between R4570 and FWHM(H$\beta_{\rm b}$) ($r = -0.55$), that belongs to the EV1 correlations firstly reported by \citet{1992ApJS...80..109B} and later confirmed in different works \citep{1996A&A...309...81W,2000ApJ...536L...5S,2001A&A...372..730V,2012AJ....143...83X}. On the contrary, \citet{2006ApJS..166..128Z} obtained only a trend between these two quantities, with $r_{\rm s} =-0.23$. Our results indicate that if an anti-correlation exists between R4570 and FWHM(H$\beta$), it is relatively weak even if significant, both for NLS1s and for the whole sample of 360 objects. However, the recent plot by \citet{2014Natur.513..210S} (see their fig.~1) approximately reproduced by means of contours in our Fig.~\ref{R4570_FW_Hb}, changed our view on this question. Indeed, it confirms what \citet{2000ApJ...536L...5S} noticed about the lack of AGNs with very broad lines and high R4570 ratios, but also shows that the anti-correlation was an apparent effect caused probably by the limited number of available data or the limited ranges of the two physical quantities involved. In the R4570--FWHM(H$\beta$) plane, the distribution of the broad line AGNs forms an horizontal sequence driven by the R4570-EW([O\,{\sc iii}]) anti-correlation, which in turn depends on the Eddington ratio, as already suggested also by \citet{2004ApJ...614..558N} and \citet{2007ApJ...654..754N}. \subsection{O\,{\sc i} and CaT properties} The origin of the O\,{\sc i}\,$\lambda$8446 was debated for long time and it is not yet completely understood \citep[see e.g.][]{1980ApJ...238...10G, 2008ApJS..174..282L}. This line seems to be produced by Ly-$\beta$ fluorescence, that excites electrons which return to the ground state through a series of transitions, producing photons at 11287, 8446 and 1302 \AA. Other possible mechanisms are collisional excitation and recombination, which can be evaluated on the basis of the presence and intensity of O\,{\sc i}\,$\lambda$7774: the ratio O\,{\sc i}\,7774/8446 should be $\sim$0.3 in case of collision \citep{1980ApJ...238...10G} and $\sim$1.1 in case of recombination \citep{2008ApJS..174..282L}. The observations show that the Ly$\beta$ fluorescence cannot explain the observed intensity of O\,{\sc i}\,$\lambda$8446 \citep{2002ApJ...572...94R,2008ApJS..174..282L,2007ApJ...663..781M} and that an additional process needs to be invoked. The continuum fluorescence, proposed by \citet{1976ApJ...207..713O}, was excluded by \citet{1980ApJ...238...10G}, \citet{2002ApJ...572...94R} and \citet{2007ApJ...663..781M} because the expected O\,{\sc i} transitions at 7254, 7990 and 13165 \AA were too weak or absent. \citet{2008ApJS..174..282L} reported to have measured the O\,{\sc i}\,$\lambda$13165 in 5 objects, even if in their figures this line is, in fact, mildly visible only in two cases. \citet{1980ApJ...238...10G} did not see O\,{\sc i}\,$\lambda$7774 in their sample of 16 Seyfert 1 galaxies. \citet{2002ApJ...572...94R} found this line only in one of their seven AGNs, while \citet{2008ApJS..174..282L} measured it in only 7 out of 23 sources and ascribed it to recombination because the O\,{\sc i}\,7774/8446 ratio was always larger than 1, but in one case. They concluded that Ly$\beta$ fluorescence contributes to O\,{\sc i}\,$\lambda$8446 for 50 per cent and more in 20 out of 23 objects and that the additional contribution is given mostly by recombination. Our data confirm that O\,{\sc i} is mostly caused by Ly$\beta$ fluorescence with an additional but apparently rather uncommon contribution given by collisional excitation. Indeed, in our sample O\,{\sc i}\,$\lambda$7254,7990, potentially visible in 307 and 161 out of 360 spectra respectively, are absent or at least too faint to be detected (O\,{\sc i}\,$\lambda$13165 is obviously not detectable in our spectral range). This excludes the continuum fluorescence as additional mechanism. O\,{\sc i}\,$\lambda$7774 was detected and measured in only 12 out of 214 spectra. The O\,{\sc i}\,7774/8446 ratios have a median value of 0.1 and range between 0.04 and 0.42, suggesting the contribution of collisional excitation. In conclusion O\,{\sc i} emission depends more on the radiation field, than on the density of the gas. CaT is a transition onto a metastable triplet produced by collisional excitation and it is known to be emitted by a high-density, optically-thick neutral gas. \citet{1988ApJ...330..751P}, \citet{2007ApJ...663..781M}, \citet{2013AstRv...8d...4M} and \citet{2015ApJS..217....3M} claimed that $\log N_{\rm H}$ between 11 and 12 and $\log U$ between $-2.5$ and $-1.5$ are required to justify CaT emission. These physical values seem to be appropriate also for O\,{\sc i}\,$\lambda$8446 \citep{2007ApJ...663..781M} but, according to \citet{2013AstRv...8d...4M} and \citet{2015ApJS..217....3M}, it is likely that O\,{\sc i} needs a higher photon flux than CaT and that it can be emitted in less extreme conditions, where CaT does not appear. We found that the intensity of Fe\,{\sc ii} is well correlated with the intensity of Ca\,{\sc ii} and slightly less with O\,{\sc i}, which could be emitted in a region with different physical properties. In fact, our models indicate that Ca\,{\sc ii} emission requires high density gas to be detected, independently on its ionization degree, and these conditions support the collisional excitation mechanism invoked to explain the frequently observed strong Fe\,{\sc ii} emission in type 1 AGNs and especially in NLS1s. High density and high ionization are mandatory for strong O\,{\sc i} emissions, while lower values of these physical parameters can explain weaker emissions. It is interesting to note that O\,{\sc i}\,$\lambda$8446 was detected and measured in 41 objects out of 67 having a spectral range including this line, while CaT only in 15 out of 52. Therefore 26 out of 41 objects with O\,{\sc i}\,$\lambda$8446 have no detection of CaT. In 11 cases, O\,{\sc i}/H$\beta_{\rm b}$ is low ($<0.15$, non reddening corrected) and these small values could require different combination of density and ionization parameter, such as lower density which would weaken CaT up to make it non detectable or hardly measurable. The other 15 objects have high values of O\,{\sc i}/H$\beta_{\rm b}$ ($\geq 0.15$) which imply, according to our models, a high density ($\log N_{\rm e}>11.5$). In these conditions, high values of CaT/H$\beta_{\rm b}$ are expected, as well. CaT can be observed in absorption due to the stellar component and this could explain the lack of detection, however NLS1s are characterized by low stellar velocity dispersion and narrow absorption lines which cannot suppress the quite strong and broad CaT emission lines. In conclusion, while on the one hand it is possible to justify why we can observe O\,{\sc i} more frequently than CaT, on the other hand the question about the apparent lack of CaT when O\,{\sc i} is strong is still open and it requires new high S/N spectra to be solved. The question of the location of the gas emitting Fe\,{\sc ii}, O\,{\sc i} and CaT is also still open and debated for long time. Our analysis shows that Fe\,{\sc ii} and H$\beta$ share a similar kinematics, even if with a large scatter. This result is in agreement with \citet{2003ApJS..145..199M}, who found a correlation between FWHM(H$\beta_{\rm b}$) and FWHM(Fe4570) in objects with FWHM(H$\beta_{\rm b}$) $< 4000$ km\,s$^{-1}$, and with \citet{2010ApJS..189...15K}, who used two Gaussians for fitting H$\beta_{\rm b}$ and showed that the Gaussian corresponding to an intermediate broad line region shares the same kinematics of Fe\,{\sc ii}. Recently \citep{2015ApJ...804..138H} confirmed this result through a reverberation-mapping analysis of a small number of AGNs and contradicted their previous findings \citep{2008ApJ...687...78H} that the kinematics of Fe\,{\sc ii} and H$\beta_{\rm b}$ are well correlated, but the FWHM(Fe\,{\sc ii}) is systematically lower than the FWHM(H$\beta_{\rm b}$). On the contrary \citet{2002ApJ...572...94R} found that Fe\,{\sc ii} and O\,{\sc i} have similar profiles, while Pa$\beta$ is much broader, suggesting that they both come from outer regions of the BLR. Similarly \citet{2016arXiv160205159M} claimed that Fe\,{\sc ii}, O\,{\sc i} and also CaT are emitted by an outer part of the BLR with respect to hydrogen \citep{2015ApJS..221...35K}. \citet{1980ApJ...238...10G} affirmed that O\,{\sc i} and H$\alpha$ profiles are very similar and excluded that O\,{\sc i} is emitted in the region of neutral hydrogen. But in fact, the FWHMs of these lines were in agreement only in 3 out of 8 cases for which the FWHM of O\,{\sc i} could be measured. We obtained a good agreement between the FWHMs of CaT and H$\beta$, adding data from \citet{1988ApJ...330..751P} and \citet{2015ApJS..217....3M}, while O\,{\sc i} FWHMs appear to be systematically lower than those of H$\beta_{\rm b}$. In other words, our results suggest that Fe\,{\sc ii}, CaT and hydrogen are emitted by BLR gas located at the same distance, while O\,{\sc i} seems to originate from a region located at a larger distance. This is in agreement with \citet{1988ApJ...330..751P} who found that the FWHMs of O\,{\sc i} and H$\beta$ correlate, but the FWHMs of O\,{\sc i} grow more slowly, and with \citet{2013AstRv...8d...4M} who said that O\,{\sc i} probably comes from regions not exactly coincident with those emitting CaT. \subsection{[O\,{\sc iii}] asymmetry and blue wings} The majority of our sample (94 per cent) shows [O\,{\sc iii}] with asymmetric profiles and no difference appears to exist between NLS1s and our BLS1s, as also found by \citet{2013MNRAS.433..622M}. \citet{2012MNRAS.427.1266V} found blue wings in 73 per cent of their sample of intermediate-type Seyfert galaxies and in 68 per cent of their sample of Seyfert 2. A slightly smaller percentage (43 per cent) was found by \citet{2016ApJ...817..108W} who claimed that outflows are common in very bright type 2 AGNs. These decreasing percentages, going from type 1 to type 2 AGNs, are in agreement with the Unified Model and suggest that the gas kinematics of the NLR is more turbulent in its inner part. From our analysis we can infer that this asymmetry is very likely caused by the presence of outflowing gas from the inner regions of the active nucleus, which interacts with the surrounding medium both by transferring kinetic energy and making it more turbulent and by reducing the EW of [O\,{\sc iii}] as its velocity increases. Indeed, in their analysis of the properties of [O\,{\sc iii}] emission line profiles, \citet{2011ApJ...737...71Z} found that the shift of the wing component correlates with the EW of the [O\,{\sc iii}] whole emission ($r_{\rm s}=0.38$, p-value $=10^{-14}$). \citet{2012ApJ...756...51L} confirmed that AGNs with blue wings in [O\,{\sc iii}] are characterized by lower values of EW and claimed that the radial motion of the clouds can decrease the covering factor of the NLR, resulting in a lower EW. The high Eddington ratio in NLS1s is in principle expected to affect also the NLR, as the resulting radiation pressure accelerates the NLR gas \citep{2005ApJ...627..721G}. However, it seems that the presence and the strength of a blue wing are not correlated with the Eddington ratio \citep{2011ApJ...739...28X,2012ApJ...756...51L}, even if high values of $L_{\rm bol}/L_{\rm Edd}$ ($>0.1$) could be associated with more prominent blue wings \citep{2013MNRAS.433..622M}. We found that the outflow velocity does not seem to depend on the AGN power and only weakly on the accretion rate and the radio luminosity. Nonetheless, we did not observe fast outflows in objects showing low luminosities and low accretion rates. Indeed, according to \citet{2011ApJ...737...71Z}, while the AGN determines the starting values of the outflow velocity, its final speed is more related to the density of the ISM: high density ISM implies more [O\,{\sc iii}] emission and decelerates the outflow more efficiently. Unexpectedly, no correlation was found between the velocity of the outflow and the electron density of the medium, in disagreement with \citet{2007ApJ...670...60X}, but in fact we measured the average density of the low-ionization gas, which is not representative of the whole medium and does not take into account possible small size and overdense regions \citep{2012MNRAS.427.1266V}. In conclusion, in our case, it is not clear whether a physical connection related to the ISM properties and in particular to its density really exists between the [O\,{\sc iii}] emission and the presence of a wing. On the other hand, high radio luminosities or accretion rates seem to be a necessary, but not sufficient condition, because the velocity of the outflowing gas and the strength of its emission must depend on the environment through which it moves. \subsection{[O\,{\sc iii}] blueshift} The radial motion of the gas seems to be responsible also for the observed blueshift of the core component of [O\,{\sc iii}]. Our analysis shows that the faster is the outflow the higher is the blueshift. In this picture the so-called blue outliers are only the objects at the extreme tail of the blueshift distribution, as already suggested by \citet{2011ApJ...737...71Z}. The continuous distribution in Fig.~\ref{v1_vS2_vs_v2} justifies the small number of observed blue outliers, since high blueshifts can be detected only in objects with very fast outflows ($v_2 - v_{\rm [S\,II]} < -400$ km~s$^{-1}$). Of course, this is a first result which must be definitely proved with spectra having higher S/N and higher spectral resolution than SDSS spectra. Fast outflows make the gas more turbulent, both considering the core and the wing components. This is more evident in blue outliers, whose FWHMs are found to increase with increasing blueshift. This trend was observed not only for [O\,{\sc iii}] lines \citep[e.g.][]{2008ApJ...680..926K}, but also for other NLR lines \citep[e.g.][]{2009ApJ...702L..42S} and for the C\,{\sc iv}\,$\lambda1549$ line \citep[e.g.][]{2013ApJ...769...30G}. High-ionization NLR lines show more frequently large shifts, as found for [O\,{\sc iii}] and [Fe\,{\sc vii}] in this work, for [Ne\,{\sc iii}] and [Ne\,{\sc v}] by \citet{2009ApJ...702L..42S} and for lines up to [Fe\,{\sc x}] by \citet{2008ApJ...680..926K}. Thus, the frequency and the strength of the shift increase with increasing ionization potential, confirming the scenario of an outflow moving through a stratified ISM ionized by the nuclear source. Moreover, the high-ionization BLR C\,{\sc iv}\,$\lambda1549$ line is found to be blueshifted by about 600 km\,s$^{-1}$ \citep{1982ApJ...263...79G} and to correlate with [O\,{\sc iii}] blueshift \citep{2002ApJ...576L...9Z}. These correlations and other similarities between [O\,{\sc iii}] and C\,{\sc iv} shift suggest that there could be a kinematic connection between BLR and NLR \citep{2002ApJ...576L...9Z,2011ApJ...737...71Z}. \citet{2013ApJ...769...30G} affirmed that an explanation for the observed blueshift could be the Rayleigh scattering associated to gas inflowing towards the SMBH. The kinematic connection between BLR and NLR gas could imply that also NLR blueshifts are due to inflow and scattering. If the radial motion of the clouds is expected to decrease the ionized gas covering factor and to reduce its emissivity, we should observe for the core component of [O\,{\sc iii}] a similar relation found between the velocity shift of the wing and the EW of the line. But this was not the case, since we did not find any correlation showing that larger blueshifts are preferentially detected in objects having weaker [O\,{\sc iii}] lines, in disagreement with \citet{2011ApJ...737...71Z}. We only observed a slight decrease of the fraction of objects with blueshifts larger than $-150$ km\,s$^{-1}$ as the EW increases. The radial motion could be caused by the accretion disk wind and therefore we should expect to find a correlation with the Eddington ratio, as found by \citet{2005ApJ...618..601A}. Indeed, \citet{2011ApJ...737...71Z} claimed that the blueshift is mainly affected by the Eddington ratio. With our data this hypothesis cannot be confirmed, because we did not find a correlation between $\Delta v$ and $L_{\rm bol}/L_{\rm Edd}$. The blueshift seems to weakly depend on the Eddington ratio, but high blueshifts are not detected at low values of this parameter. Interestingly, even the blue outliers appear not to depend on the Eddington ratio, but rather on the BH mass and the luminosity of the active nucleus. However, the Eddington ratio is calculated by means of these two quantities, therefore, if the correlations of both continuum luminosity and BH mass with the velocity shift of [O\,{\sc iii}] are similar, the dependency on the Eddington ratio could disappear. Our blue outliers show Eddington ratios distributed within the range of values of the whole sample and they are not those AGNs with the highest values of this quantity (see top and middle panels of Fig.~\ref{hist_eddington}). However, the frequency of blue outliers increases with increasing Eddington ratio (see bottom panel of Fig.~\ref{hist_eddington}). Indeed, if it is true that blue outliers are often characterized by high $L_{\rm bol}/L_{\rm Edd}$ ratios \citep{2003MNRAS.345.1133M,2008ApJ...680..926K}, not all high Eddington ratio AGNs with narrow H$\beta_{\rm b}$ are blue outliers \citep{2005AJ....130..381B, 2005ApJ...618..601A}. \begin{figure} \includegraphics[width=\columnwidth]{fig31.eps} % \caption{Comparison between the distribution of the Eddington ratios for our whole sample (top panel), the distribution of the same ratio for our blue outliers (middle panel, solid line) and for the whole sample of blue outliers (middle panel, dashed line). The bottom panel shows the frequency of blue outliers.} \label{hist_eddington} \end{figure} Finally, the blueshift of [O\,{\sc iii}] core component does not seem to be affected by the radio emission, as already found for the wing component. We believe that a powerful radio emission is mandatory to significantly affect the NLR kinematics. Since only 11 NLS1s in our sample are radio-loud, the lack of a correlation with radio properties is expected. Indeed, \citet{2009ApJ...702L..42S} found a correlation between the [Ne\,{\sc iii}] linewidth and the radio luminosity for high luminosity objects and claimed that the interaction of the radio jet with the ISM could explain the observed kinematics. Furthermore, \citet{2015arXiv150605800B} found that blue outliers are more frequent in radio-loud sources likely because of an interaction between the NLR and the relativistic jets.

1607.03438

1607

1607.08036_arXiv.txt

Robust knowledge of molecular gas mass is critical for understanding star formation in galaxies. The \htwo\ molecule does not emit efficiently in the cold interstellar medium, hence the molecular gas content of galaxies is typically inferred using indirect tracers. At low metallicity and in other extreme environments, these tracers can be subject to substantial biases. We present a new method of estimating total molecular gas mass in galaxies directly from pure mid-infrared rotational \htwo\ emission. By assuming a power-law distribution of \htwo\ rotational temperatures, we can accurately model \htwo\ excitation and reliably obtain warm ($T\!\gtrsim\!100$\,K) \htwo\ gas masses by varying only the power law's slope. With sensitivities typical of Spitzer/IRS, we are able to directly probe the \htwo\ content via rotational emission down to $\sim80$\,K, accounting for $\sim15\%$ of the total molecular gas mass in a galaxy. By extrapolating the fitted power law temperature distributions to a calibrated \emph{single} lower cutoff temperature, the model also recovers the total molecular content within a factor of $\sim$2.2 in a diverse sample of galaxies, and a subset of broken power law models performs similarly well. In ULIRGs, the fraction of warm \htwo\ gas rises with dust temperature, with some dependency on $\alpha_\mathrm{CO}$. In a sample of five low metallicity galaxies ranging down to $\twelveoh=7.8$, the model yields molecular masses up to $\sim 100\times$ larger than implied by CO, in good agreement with other methods based on dust mass and star formation depletion timescale. This technique offers real promise for assessing molecular content in the early universe where CO and dust-based methods may fail.

} By a factor of more than $10^4$, \htwo\ is the most abundant molecule in the universe, found in diverse environments ranging from planet atmospheres to quasars hosts \citep{Hanel79, Hanel81, Walter03, Genzel14}. It was the first neutral molecule to form in the Universe and hence dominated the cooling of pristine gas at early times \citep{Lepp02}. Stars form principally from molecular clouds, and most physical prescriptions for stellar formation rate (SFR) are therefore directly linked to the surface density of \htwo\ gas \citep{Kennicutt98, Bigiel08}. To understand the star formation processes and how it varies and has evolved over the star-forming history of the Universe, it is necessary to accurately measure the mass and distribution of \htwo. Despite its high abundance, \htwo\ can be difficult to study directly. It possess no permanent dipole moment, which makes it a weak rotational emitter. In addition, the upper energy level for the lowest permitted rotational (quadrupole) transition is $E/k=510$ K above ground \citep{Dabrowski84, Dishoeck86}. Hence, the \htwo\ gas that comprises the bulk of the molecular interstellar medium (ISM) is commonly believed to be too cold to be visible. And yet, despite these observational disadvantages, the advent of the Infrared Space Observatory (ISO) and in particular Spitzer's Infrared Spectrograph \citep[IRS,][]{Houck04} revealed pure rotational \htwo\ emission from a rich variety of extragalactic sources, including normal star forming galaxies \citep{Roussel07}, ultraluminous and luminous infrared galaxies \citep[U/LIRGs][]{Lutz03, Pereira10, Veilleux09, Stierwalt14}, galaxy mergers \citep{Appleton06}, radio-loud AGN \citep{Ocana10}, UV-selected galaxies \citep{ODowd09}, quasar hosts \citep{Evans01}, cooling-flow cluster systems \citep{Egami06}, sources with extreme shock-dominated energetics \citep{Ogle10}, and even star-forming sources at redshifts $\gtrsim 2$ \citep{Ogle12}. In the absence of direct measurements of \htwo, the lower rotational line transitions of CO (the next most abundant molecule) are generally used as a molecular gas tracer. To convert the measured integrated intensities of $^{12}$CO to \htwo\ column density a conversion factor, $\rm{\alpha_{CO}}$, is needed, typically calibrated against virial mass estimates in presumed self-gravitating clouds \citep{Solomon87, Scoville87, Strong96, Abdo10}. However recent evidence indicates that the value of $\rm{\alpha_{CO}}$ varies substantially, both within galaxies and at different epochs \citep{Genzel12, Sandstrom13}. In molecular clouds in our Galaxy, the observed $\gamma$-ray flux arising from the cosmic ray interactions with \htwo\ can be used to recover \htwo\ gas mass accurately \citep{Bhat85, Bloemen84, Bloemen86}, but this method requires information on the cosmic ray distribution not available in other galaxies. Another method to assess \htwo\ content is to use dust as a tracer, with a presumed or recovered dust-to-gas (DGR) ratio \citep[e.g.][]{Sandstrom13, Remy14}. Among other biases, both the CO and dust-based methods lose reliability at low metallicity. Due to declining dust opacity and less effective self-shielding than \htwo, the CO abundance drops rapidly with decreasing metallicity \citep{Wolfire10, Bolatto13}. Variations in the radiation field can also lead to selective CO destruction \citep{Hudson71, Bolatto13}. At the lowest metallicities, DGR itself begins to scale non-linearly with the metal content of the gas \citep{Herrera12, Fisher14, Remy14}. Another method of inferring molecular gas content is to couple measured star formation rates with an assumed constant \htwo\ depletion timescale ($\tau_{dep}$), equivalent to a constant star formation efficiency (SFE) in galaxies \citep{Schruba12}. None of these methods provide \emph{direct} tracers of the \htwo\ reservoirs, and the relevant indirect tracers all rely on physical assumptions (e.g. constant or measurable values of $\rm{\alpha_{CO}}$, DGR, or $\tau_{dep}$) which are unlikely to be valid in all environments. Here we challenge the long-stated assumption that \htwo\ rotational emission is a poor tracer of the total molecular content in galaxies. Our understanding of the structure of the molecular material in galaxies is undergoing significant revision. While CO traces the coldest component of the molecular ISM, half or more (and sometimes much more) of the molecular gas in galaxies is in a warmer, CO-dark state, where \htwo\ persists \citep{Field66, Tielens05, Draine11, Wolfire10, Pineda13, Velusamy14}. What this means is that the common wisdom that all molecular gas is found at temperatures $T=10-20$\,K is incorrect. With molecular material existing in quantity at excitation temperatures 50 - 100\,K, it becomes possible to use the rotational emission spectrum of the Universe's dominant molecule to directly assess a substantial portion of the total molecular content in galaxies. We introduce here a simple continuous temperature model which enables direct use of the \htwo\ rotational emission from galaxies to recover their total molecular gas content. While several assumptions are required to make such a model possible, the resulting biases are expected to be completely distinct from those of the indirect gas tracers. The paper is laid out as follows. We describe the archival sample in \S\,2. A rotational temperature distribution model is presented in \S\,3, and in \S\,4 we describe the method and its calibration procedure. Section 5 presents results, discussions, applications and future prospects of our model. We summarize our conclusions in \S\,6.

\subsection{What are the typical molecular gas temperatures in galaxies?} UV pumping sets the level populations of upper level states, J $\geqslant$ 3, in particular at low column densities of \htwo. As a result, these states may result in rotational temperatures well in excess of their kinetic temperatures. For our purposes, whatever combination of collisional (including shocks) and UV pumping determines the level populations and hence average the \htwo\ excitation, we assume a single power law distribution of rotational temperature describes the ensemble of molecules (although see \S\,\ref{sec:broken-power-law} for an alternative analysis employing broken power-laws). Since the mass is dominated by gas with the lowest rotational temperatures and hence lowest excitations, it is instructive to compare the derived lower temperature cutoffs (\S 4.2.2) with various theoretical models and measurements of temperature in molecular gas. The average lower extrapolation cutoff, $T_\ell^\star=49$\,K, is comparatively higher than what is typically assumed for molecular regions ($\sim$10--30\,K). And yet, this high lower cutoff is essential in describing the portion of molecular material following a single power-law distribution of temperatures. For example, if we forced the extrapolation temperature to a lower value of 20\,K while retaining the average power law index $n=4.8$ (\S\,4.1.2), the estimated molecular gas mass would be on average $\sim30\times$ higher than the mass measured using $\rm{L_{CO}}$. The ratio of warm diffuse to cold dense molecular gas in galaxies is a key step in understanding the \htwo\ temperature distribution. Assuming an incident radiation field G$_{0}$ = 10 for a cloud with mass 10$^{5}$--10$^{6}$ M$_{\odot}$, \citet{Wolfire10} estimated that molecules in the outer diffuse region ($A_{V} <1$) obtain temperatures in the range 50--80\,K, with stronger incident radiation fields leading to even higher temperatures. Due to the differences in self shielding from dissociating radiation, molecular clouds have an outer layer of varying thickness which contains molecular hydrogen, but little or no CO. This gas cannot be traced through CO transitions and is hence called \emph{dark molecular gas} \citep{Wolfire10}. This dark gas can account for a significant fraction (24--55\%) of the total molecular gas mass in galaxies \citep{Pineda13, Wolfire10, Smith14, Planck11}. CO-dark and diffuse molecular gas is heated to higher temperatures than dense CO-emitting gas in molecular cloud cores. Indeed, individual molecular cloud simulations find average mass-weighted \htwo\ temperatures of $\sim$45\,K for Milky-Way average cloud masses and radiation (\citet{Glover12a}, Glover, priv comm.). Galaxy-scale hydrodynamic simulations with a full molecular chemistry network \citep{Smith14} also show that \htwo\ gas in the CO-emitting regions is markedly biased to the low temperature end of the full temperature distribution (Glover, priv comm.). Also, in diffuse and translucent molecular clouds in the Galaxy, \citet{Ingalls11} demonstrated that additional sources of heating are required to explain the observed \htwo\ and atomic cooling-line power. Our extrapolated power law model traces both warm and cold molecular gas mass in galaxies. The common assumption that the bulk of molecular gas is found at rotational temperatures of $\sim$10--30 K may be true only in the CO emitting core regions of the molecular cloud. The average mass-weighted molecular gas temperature can be much higher when the full range of emitting environments are considered. It is therefore perhaps not surprising to find typical molecular gas temperatures of $\sim50$\,K in galaxies, similar to our model derived values. It must be noted, however, that to calculate the total molecular gas mass, we have extrapolated using a single power law index. The possibility of a broken or non-power-law temperature distribution for the $\sim$85\% of \htwo\ gas at temperatures lower than our sensitivity temperature cannot be excluded (and indeed in warm gas excess sources, will be required to explain ongoing star formation). In the following section, we consider the effect of broken power law models on the dominant \htwo\ temperatures. \subsection{Broken power law models} \label{sec:broken-power-law} In the previous subsection we discussed the results of applying a single power law model for the entire temperature distribution of \htwo, and found that a single lower cutoff temperature of $T_{\ell}^{\star}$ = 49\,K permits recovering the entire molecular mass. Yet as seen in \S\,\ref{sec:determ-model-extr}, with the typical performance of the Spitzer/IRS instrument, we are sensitive only to the $\sim$15\% of \htwo\ gas with temperature $T\gtrsim T_s=81$\,K. While a single power law temperature distribution, varying only the slope $n$, provides very high quality model fits to the \htwo\ excitation diagram, at temperatures below $T_s$, in principle \emph{any} distribution form would provide an equally valid description. Chemo-hydrodynamic models with \htwo\ formation and excitation do result in a broad distribution of \htwo\ temperatures (Glover \& Smith, in prep.). But shocks elevate \htwo\ to higher rotational temperatures \citep[e.g.][]{Neufeld08}, and are not well modeled, so as yet there exists no \emph{a priori} expectation for the detailed 10--3000\,K temperature distribution of \htwo\ molecules. Any presumed distribution, however, must 1) reproduce the \htwo\ excitation diagrams as well as a single power law does, and 2) permit recovering the CO-predicted \htwo\ mass within a reasonable range of temperatures. A \emph{broken power-law} provides a general framework for exploring changes in the low-temperature form of the distribution, and has been invoked by \citet{Pereira14} to jointly model \htwo\ and CO excitation. Here we consider the changes resulting from adopting a broken power law model, with a fixed break temperature, $T_{b}$, and slope below the break, $n_1$, studying the resulting lower cutoff temperatures, as well as the ability of these models to satisfy these two criteria. Holding the upper temperature $T_u$ constant as before, we evaluated a grid of broken power law models with fixed pre-break slope $n_1$ (from 0--3) and break temperature $T_{b}$ (from 70--290\,K). For each such model, we fitted the observed MIR \htwo\ rotational line ratios in excitation diagrams in the same training sample of 34 galaxies considered in \S\,\ref{sec:model-extr-lower}. As before, we then extrapolate the broken power law to recover the total \htwo\ mass, as indicated by CO intensity, evaluating the associated extrapolation cutoff temperature $T_{\ell,b}$. We define $T_{\ell,b}^{\star}$ as the average value of $T_{\ell,b}$ across the training sample, analogous to $T_{\ell}^{\star}$ for a single power law (\S\,\ref{sec:model-extr-lower}). \begin{figure} \centering \includegraphics*[width=0.5\textwidth]{fig9.eps} \caption{The lower extrapolation temperature for broken power-law models, $T_{\ell,b}^{\star}$, as a function of the model break temperature, $T_{b}$ for different values of lower power law indices $n_1$ = 0 (black), 1 (red), 2 (green), and 3 (blue). The solid horizontal line denotes a lower temperature physical cut off of 5\,K. The dotted line in the figure shows the range where the model poorly fits the MIR $\htwo$ rotational lines, with median $\Delta\chi^{2} > 2.3$. At break temperatures above $\sim$120\,K, for all n$_{1}$ (0--3), the $\htwo$ rotational lines are poorly fit.} \label{fig:fig9} \end{figure} Figure \ref{fig:fig9} shows the variation of the training sample's average extrapolated lower cutoff temperature $T_{\ell,b}^{\star}$ as a function of break temperature $T_{b}$ for differing values of below-the-break slope. Those combinations of ($n_1$,$T_b$) for which the excitation diagrams are reproduced as well as a single power law model ($\Delta\chi^{2} < 2.3$ relative to the single power law model fit; see \S\,\ref{sec:determ-model-extr}) for at least 50\% of the training sample are shown in solid lines. In this allowable range of $n_1$ and $T_b$, the dispersion of the ratio $M(\htwo,\mathrm{model})/M(\htwo,\mathrm{CO})$ of \htwo\ mass recovered from the broken power-law model using a single extrapolation $T_{\ell,b}^{\star}$, relative to that implied by CO ranges only from 0.28--0.34 dex. At higher temperatures the excitation diagram fit quality degrades rapidly (dotted lines). As the break temperature approaches the average slope for a single power law model, $T_b=T_\ell^\star=49$\,K, the broken power law becomes equivalent to our single power law truncated at this temperature, independent of slope $n_1$. At slopes approaching the single-power law average $n=4.8$ (see Fig.~\ref{fig:fdspl}), lower extrapolated temperatures diverge less and less from the value obtained using a single power law, since the shape of the temperature distribution is then relatively unchanged. Low slopes $n_1$ require very low extrapolation temperatures, since the cumulative mass of \htwo\ rises less rapidly at lower temperatures compared to the single power law. In fact, models with a flat slope ($n_1=0$) fail to recover the CO-derived \htwo\ mass within a reasonable lower \htwo\ temperature floor of $\sim5$\,K. Figure \ref{fig:fig9} demonstrates that the single lower cutoff temperature to which an \htwo\ model must be extrapolated depends sensitively on the detailed shape of the distribution at low temperature. For reasonable slopes, the dominant mass-weighted temperature can range from 10--50\,K. All models which produce satisfactory fits of the excitation diagrams and have physically reasonable extrapolation temperatures have nearly the same predictive capability for the \emph{total} \htwo\ mass. While optical depth effects introduce some difficulties (e.g. for low-J $^{12}$CO), ideally a family of profiles for the full temperature distribution of molecular gas in galaxies could be developed, based on tracers with sensitivity to different parts of the temperature range matched to chemo-hydrodynamic models of molecular gas across all densities. Since the single power law model provides excellent fits to \htwo\ rotational emission line ratios with the minimum of free parameters, we retain this simplified model for the remainder of this work. It should, however, be kept in mind that broken power law models with flatter slopes below $T\sim$100--120\,K can produce similar quality fits and yield equivalently reliable model-based \htwo\ masses, as long as the low temperature slope $n_1$ does not vary substantially from galaxy to galaxy. \subsection{Estimating total $M_{H_{2}}$} \begin{figure*} \centering \includegraphics*[width=1.00\textwidth]{fig10.eps} \caption{The model extrapolated molecular gas mass, assuming a single lower cutoff temperature of 49\,K vs the total molecular gas mass obtained from L$\rm{_{CO}}$ measurements using $\rm{\alpha_{CO,Gal}}$. Different symbols represent different galactic systems as indicated. The solid line is one to one correspondence while the dashed and dotted lines are $2\times$ and $3\times$ the value respectively.} \label{fig:49co} \end{figure*} By fitting a continuous temperature distribution to the MIR \htwo\ rotational lines and calibrating an extrapolating temperature, this model can be used to calculate the total molecular mass in galaxies directly from \htwo\ rotational emission. While this method does require a reliable source of known molecular masses for calibration (introducing secondary dependence on, e.g. $\alpha_{CO}$), once calibrated it is independent of any indirect tracer like CO, DGR, or assumptions about star formation depletion timescales. The biases inherent in this method (arising, for example, from the assumption of a single and smooth power law temperature distribution), are therefore expected to be relatively distinct from those of the aforementioned estimators. In this section we test the model's capability to estimate molecular gas mass in different types of galaxies. Adopting a fixed model lower extrapolation temperature $T_{l}^{\star} = 49$ K the total \htwo\ gas mass is calculated by extrapolating the fitted model. The total \htwo\ gas mass derived by our model is shown in Figure \ref{fig:49co}. It compares very well with $\rm{L_{CO}}$ along with $\rm{\alpha_{CO,Gal}}$ based estimates of gas mass. The scatter in our model is about 0.31 dex (factor of 2) for the SINGS sample and increases to 0.34 dex (factor of 2.2) for the complete sample, including U/LIRGs and radio galaxies. Some of this scatter no doubt arises from ignorance of the true $\rm{\alpha_{CO}}$ values in these systems. Galaxies with warm gas excess and a high sensitivity temperature $T_{s}$ (\S\,\ref{sec:determ-model-extr}), require similar values of extrapolated $T_{\ell}$ as the training sample. The warm molecular gas traced by MIR rotational lines may be completely isolated from the cooler gas in these galaxies. Extrapolating to a lower temperature of 49 K yields molecular gas masses which agree with CO-derived values, assuming $\rm{\alpha_{CO}=\alpha_{CO,Gal}}$. A possibility of enhanced CO excitation due to high molecular gas temperature with a corresponding low $\rm{\alpha_{CO}}$ in such warm galaxies cannot be ruled out (see also \S\,\ref{sec:model-deriv-molec}). Taken together, a tight correlation between \htwo-derived and CO-based mass estimates is found, spanning seven orders of magnitude in mass scale and across a wide range of galaxy types. The mass calculated with a continuous temperature distribution model, extrapolated to a single \emph{fixed} cutoff temperature of 49\,K can provide an independent measurement of total molecular gas mass in galaxies, good to within a factor of 2.2 (0.34 dex). A dependence on $\alpha_{\mathrm{CO}}$ does arise indirectly through our CO-based mass estimates in the training sample (\S\,\ref{sec:model-extr-lower}), but the recovered dispersion is comparable to uncertainties in the $\rm{\alpha_{CO}}$ conversion factor itself, even among normal galaxies \citep{Bolatto13}, as well as methods using DGR \citep[e.g.][]{Sandstrom13}. \subsection{Model derived molecular gas mass in ULIRGs, LIRGs and radio galaxies} \label{sec:model-deriv-molec} Many local ULIRGs are recent ongoing galaxy mergers. In the merging process a large amount of gas in the spiral disk is driven to the central nuclear region, increasing the gas temperature. The increase in temperature and turbulence increases the CO linewidth, resulting in a high value of $\rm{L_{CO}}$ for a given molecular gas mass. $\rm{\alpha_{CO,Gal}}$ therefore gives an overestimate of \htwo\ gas masses in these galaxies \citep{Downes93, Bryant99}. Moreover, $\rm{\alpha_{CO,Gal}}$ can yield molecular gas masses greater than the observed dynamical mass \citep{Solomon97}. To avoid this, a lower value of conversion factor is suggested for ULIRGs and other merger systems --- $\rm{\alpha_{CO}}$ = 0.8 M$_\odot$(K km s$^{-1}$ pc$^{2})^{-1}$, 5.5$\times$ lower than the standard Galactic value \citep{Downes98}. However, by considering the high-J CO ladder, some studies have suggested that even for ULIRGs $\rm{\alpha_{CO,Gal}}$ values are possible \citep{Papadopoulos12}. Some \htwo\ emission may lie outside photo-dissociation and star forming regions in ULIRGs \citep{Zakamska10}. This \htwo\ may reside in CO-dark gas, so that applying a low $\rm{\alpha_{CO}}$ value in ULIRGs may yield an underestimate \htwo\ mass. In radio galaxies molecular gas can be predominantly heated by shocks through powerful jets. The molecular gas clouds may be affected by turbulence, and not gravitationally bound, defying the use of standard $\rm{\alpha_{CO,Gal}}$. \citet{Ogle14} using DGR in NGC\thinspace4258, a low luminosity AGN (LLAGN) harboring a jet along the disk, derived gas mass of about 10$^{8}$ M$_\odot$, an order of magnitude lower compared to the standard method of using $\rm{\alpha_{CO,Gal}}$. The molecular gas mass could be overestimated when used $\rm{\alpha_{CO,Gal}}$ in radio galaxies, which harbor long collimated powerful jets. In applying a power law model to the sample of ULIRGs, LIRGs and radio galaxies, the nominal model extrapolation temperature $T_\ell^\star=49$\,K is used to calculate the total \htwo\ gas masses, listed in column 3 of Table \ref{table:uli}. The $T_{\ell}$ in column 4 of Table \ref{table:uli} is the required extrapolated temperature to match the cold molecular gas mass measured from the CO line intensity. In radio galaxies 3c\thinspace424, 3c\thinspace433, cen\thinspace A, and 3c\thinspace236 the estimated \htwo\ gas mass using the power law model after extrapolation to $T_{\ell}^{\star}$ = 49 K, is higher when compared to the CO luminosity derived values using $\rm{\alpha_{CO,Gal}}$. However, when accounted for the intrinsic variation in $\rm{\alpha_{CO,Gal}}$ and $T_{\ell}^{\star}$ the \htwo\ gas masses are in agreement with each other except in 3c\thinspace424, where the difference in masses is more than 10$\times$. Using $\rm{\alpha_{CO}}$ = 0.8 M$_\odot$(K km s$^{-1}$ pc$^{2})^{-1}$, a factor of 5.5$\times$ lower than the standard Galactic value, we derive a modified extrapolation temperature $T'_{\ell}$, which is required to match the lowered gas mass. Since the model mass rises rapidly to lower temperatures, $T'_{\ell} > T_{\ell}$. The $T'_{\ell}$ values are listed in column 7 of Table \ref{table:uli}. On average for ULIRGs, $T'_{\ell}=80\pm13$\,K. Figure~\ref{fig:ulihist} shows the temperature distribution of $T_{\ell}$ (adopting $\rm{\alpha_{CO,Gal}}$) and $T'_{\ell}$ (adopting $\rm{\alpha_{CO,Gal}}/5.5$) in ULIRGs. The lower temperature cutoff in ULIRGs and radio galaxies (except 3c424) is very similar to the normal galaxy training sample when $\rm{\alpha_{CO,Gal}}$ is adopted, but much higher with the reduced molecular mass of $\alpha_\textrm{CO,ULIRG}$. It is of interest that when $\alpha_\textrm{CO,Gal}$ is used with the nominal calibration cutoff temperature $T_\ell^\star$, the ULIRG sample in Fig.~\ref{fig:49co} does not exhibit any particular bias. Either $\alpha_\textrm{CO}$ and $T_\ell$ are similar to their normal Galactic values in these systems, or they have reduced $\alpha_\textrm{CO}$ and a higher \htwo\ temperature floor. This may indicate that the same physical processes that lead to reduced $\alpha_\textrm{CO}$ in highly active system, including increased ISM pressure and radiation density, globally elevate the gas temperature. Since ULIRGs could indeed have uniformly elevated molecular gas temperatures, this degeneracy between $\rm{\alpha_{CO}}$ decrease and $T_\ell$ increase leads to a systematic uncertainty in the total \htwo\ gas mass identical in form to that obtained directly from mass estimates based on $\rm{L_{CO}}$. A suggested prescription which side-steps this ambiguity, in applying this model to systems with non-Galactic $\rm{\alpha_{CO}}$ is to calculate the total gas mass using the the nominal $T_\ell^\star=49$\,K, and scale it by $\alpha_\textrm{CO}/\alpha_\textrm{CO,Gal}$, for the preferred $\alpha_\textrm{CO}$. As can be seen in Fig.~\ref{fig:49co}, which adopts a uniform $\rm{\alpha_{CO,Gal}}$, the molecular mass in ULIRGs is well recovered by this procedure. \begin{figure} \centering \includegraphics*[width=0.5\textwidth]{fig11.eps} \caption{The distribution of model extrapolated lower temperatures in ULIRG and radio galaxies when the \htwo\ gas mass is evaluated using the Galactic conversion factor $\rm{\alpha_{CO,Gal}}$ (above, red), and when adopting $\rm{\alpha_{CO,Gal}/5.5}$, as generally accepted for ULIRGs (below, blue). The mean lower temperature cutoff is 50\,K and 80\,K when $\rm{\alpha_{CO,Gal}}$ and $\rm{\alpha_{CO}}=\rm{\alpha_{CO,Gal}}/5.5$ are used, respectively. The $T_\ell$ distribution for the SINGS normal galaxy sample is shown above, for comparison.} \label{fig:ulihist} \end{figure} \subsection{Effect of dust temperature on the warm \htwo\ fraction} \label{sec:effect-dust-temp} Since typical sensitivity temperatures are $T_s\sim80$\,K (see \S\,\ref{sec:determ-model-extr}), we can directly calculate the warm \htwo\ mass above $\sim100$\,K without extrapoloation, and compare it to the total mass as recovered by the calibrated extrapolation of \S\,\ref{sec:model-extr-lower}. Given an estimate for the power law index, $n$, and lower cut off temperature, $T_{\ell}$, of the power law distribution, we can calculate the fraction of molecular gas mass at temperatures above 100\,K as \begin{equation} \frac{M(>100\,\mathrm{K})}{M_{\mathrm{total}}} = \frac{M(>100\,\mathrm{K})}{M(H_{2},CO)} = \frac{\int_{100\mathrm{K}}^{T_{u}} T^{-n} dT}{\int_{T_{\ell}}^{T_{u}} T^{-n} dT}, \end{equation} where M$_{\mathrm{total}}$ is the total molecular gas mass estimated from the CO line intensity. Assuming $100\,\mathrm{K} \ll T_{u}$ and $T_{\ell} \ll T_{u}$ we find \begin{equation} \frac{M(>100\,\mathrm{K})}{M_{\mathrm{total}}} \approx \left(\frac{100\,\mathrm{K}}{T_{\ell}}\right)^{1-n}. \label{eqn:warmfrac} \end{equation} Table \ref{table:modelvalue} lists the calculated mass fraction $\rm{M(> 100 K)/M_{total}}$ for each galaxy. Columns 6 and 8 of Table \ref{table:uli} are calculated using $\rm{\alpha_{CO,Gal}}$ and $\rm{(1/5.5)\times\alpha_{CO,Gal}}$ as generally assumed for ULIRGs, respectively. At lower $\rm{\alpha_{CO}}$, the warm gas mass fraction is higher. \begin{figure*} \centering \includegraphics*[width=1.0\textwidth]{fig12.eps} \caption{The fraction of warm \htwo\ gas mass ($T>100$\,K) versus the dust color temperature, $\rm{\nu f_{\nu 70 \micron}/\nu f_{\nu 160 \micron}}$, obtained from PACS. At left a consistent Galactic $\rm{\alpha_{CO}}$ is adopted, and little trend is seen. At right, a reduced $\rm{\alpha_{CO}}$ is applied to ULIRGs and QSO, and the dust-derived central $\rm{\alpha_{CO}}$ values of \citet{Sandstrom13} are used. Galaxies with warmer dust color temperatures have high warm molecular gas mass fraction.} \label{fig:wts} \end{figure*} Figure \ref{fig:wts} shows the fraction of warm molecular gas mass, $\rm{M(> 100 K)/M_{total}}$, as a function of far infrared (FIR) dust color temperature, $\rm{\nu f_{\nu 70 \micron}/\nu f_{\nu 160 \micron}}$. The warm gas fraction obtained using $\rm{\alpha_{CO,Gal}}$ ranges from 2--30\%, and exhibits little correlation among different galaxy types with dust color temperature. The ULIRGs and normal star forming galaxies show similar warm gas fractions, though ULIRGs have warmer dust color temperatures, and LINERs and Seyferts have somewhat higher warm gas mass fractions than normal star forming galaxies on average. In contrast, using the available dust-derived central $\rm{\alpha_{CO}}$ estimates of \citet{Sandstrom13} for normal galaxies and a reduced value $\frac{1}{5.5}\times\rm{\alpha_{CO,Gal}}$ for ULIRGs and QSO's, however, leads to a strong correlation, with warmer dust implying an increasing warm \htwo\ fraction. The average warm gas fraction for ULIRGs and QSO is then $\sim$45\%, significantly above that of normal galaxies, and on the same increasing trend with dust temperature. Depending on which prescription for total gas mass is correct, this could indicate a dependence of $\rm{\alpha_{CO}}$ on temperature. When considering a broken power law model, the fraction M($>$100 K)/M$_{total}$ is, as expected, very similar to that of a single power law for $T_{b} \leq$ 100 K. For $T_{b} >$ 120\,K irrespective of low-temperature slope $n_{1}$, the model yields a poor fit to the excitation diagrams (see \S\,\ref{sec:broken-power-law}). At these intermediate break temperatures, the warm mass fraction M($>$100 K)/M$_{total}$ changes by at most $\sim$10--20\% compared to the case of a single power law. \subsection{Molecular gas in low metallicity galaxies} As traced by their CO emission, many low metallicity dwarfs appear to have vanishingly low molecular gas content, but retain high star formation rates. That is, they are strong outliers on the Schmidt-Kennicutt relation \citep{Galametz09, Schruba12}. This discrepancy implies either that in dwarf galaxies star formation efficiencies are higher compared to normal spirals, or they host large molecular gas reservoirs than is traced by the CO emission \citep{Schruba12, Glover12b}. It is possible that a significant fraction of \htwo\ exists outside the CO region, where the carbon is in C$^{+}$ (ionized) or C$^{0}$ (neutral) states. Since \htwo\ can self-shield from UV photons in regions where CO is photodissociated \citep{Wolfire10}, at low metallicity, not only is the CO abundance reduced, but as dust opacity is reduced and ionized regions become hotter and more porous, $\alpha_{\rm{CO, Gal}}$ can severely underestimate the molecular gas mass. Considerable effort has been invested in detecting and interpreting CO emission at metallicity 50 times lower than the solar metallicity, \twelveoh$\sim7.0$, to assess the molecular gas content \citep{Leroy11, Schruba12, Cormier14, Remy14}. Dust emission can be used to estimate the molecular gas mass in ISM, assuming a constant DGR however, it is essential to know the change in DGR with metallicity. At very low metallicities, \twelveoh$\le8$, DGR appears to scale non-linearly with metallicity \citep{Herrera12, Remy14}. Our direct detection of \htwo\ gas mass through \htwo\ rotational lines is independent of any indirect tracers, which are affected by changing metallicity and local radiation effects. Applying our model in low metallicity galaxies should yield an estimate of molecular gas mass without the same inherent biases introduced by these dust and CO abundance variations. In this section we estimate the \htwo\ gas masses through our power law model in a low metallicity galaxy sample selected to have detected \htwo\ rotational emission, faint CO detection, and (where available) estimates of dust mass. We then compare these \htwo -based estimates to models and other methods which attempt to control for the biases introduced at low metallicity. The low metallicity galaxies were selected on availability of MIR \htwo\ rotational lines to have atleast three rotational lines including S(0) or S(1) line fluxes along with CO derived molecular mass estimates. \subsubsection{Metallicity estimation} To study the variation of \htwo\ gas mass from CO derived measurements over the metallicity range, it is essential to estimate the metallicity of galaxies. The metallicities were determined applying the direct $T_{e}$ method. CGCG\thinspace007-025 is the lowest metallicity galaxy in our dwarf sample with the value of \twelveoh = 7.77 \citep{Izotov07}. \citet{Guseva12} estimated the value of \twelveoh\ in the two H II regions, Haro\thinspace11B and Haro\thinspace11C as 8.1 and 8.33, respectively hence, we adopt the average value for Haro\thinspace11. The metallicity value, \twelveoh, for NGC\thinspace6822 is 8.2 \citep{Israel97}. No literature value exist for the oxygen gas phase abundance for the specific region of Hubble V in NGC\thinspace6822, mapped by the IRS-Spitzer. \citet{Peimbert05} estimated \twelveoh = 8.42$\pm$0.06 for Hubble V, which is inconsistent with the previous value of 8.2. For selected SINGS galaxies, the metallicity values in the circumnuclear regions, which are approximately the size of our \htwo\ line flux extracted regions, are estimated by averaging the theoretical (KK04) and an empirical metallicity calibration (PTO5) as recommended by \citet{Moustakas10}. \subsubsection{Cold molecular gas from CO line emission} Although the CO abundance drops super-lineraly with decreasing metallicity, it is detected in the low metallicity sample, and as the most common molecular tracer, can be compared directly to our \htwo\- based method. We adopted the literature values for $^{12}$CO(1--0) line intensities for CGCG\thinspace007-025 and N66, while for Haro\thinspace11, UM\thinspace311 and Hubble\thinspace V region $^{12}$CO(3--2) line intensities were scaled using the relation $\rm{I_{CO(3-2)}/I_{CO(1-0)}}$ = 0.60 (in temperature units) due to unavailability of $^{12}$CO(1--0) line intensities. Calculating $\rm{L\arcmin_{CO}}$, (area integrated luminosity) the molecular gas masses are estimated using $\rm{\alpha_{CO,Gal}}$ and were further scaled using the 8 $\micron$ map, to account for the difference in the extracted IRS spectrum and the CO beam regions for each galaxy. The molecular gas masses are listed in Table \ref{table:lowz}. \subsubsection{Molecular gas from dust emission} An alternative method for estimating molecular gas mass makes use of dust emission together with assumption of dust opacity and grain size distribution to calculate a total mass, scaling dust mass to the total gas mass using a presumed or modeled dust-to-gas ratio, and removing the measured atomic mass from the region. \citet{Leroy07} estimated \htwo\ gas surface density using dust emission from FIR map in N66 region of SMC. They derived $\rm{\alpha_{CO}}$ to be about 27$\times\rm{\alpha_{CO,Gal}}$. For Haro\thinspace11, UM\thinspace311 and Hubble\thinspace V using metallicity-DGR relation from \citet{Sandstrom13} and with the known dust mass, we estimated the total gas mass and after subtracting the atomic gas content subsequently calculated the molecular gas mass. However, we find a negative value for the molecular gas for UM311, suggesting the ISM mass is mainly dominated by the atomic gas. The \htwo\ gas masses calculated from the dust emission are given in Table \ref{table:lowz}. \subsubsection{Molecular gas mass using our model} The \htwo\ line flux extraction was performed for the similar region in SMC-N66, where the CO emission was measured by \citet{Rubio96}. The cubes were prepared using CUBISM \citep{Smith07b} (Jameson, K. et al. in prep) and the \htwo\ line fluxes were estimated using the PAHFIT, a MIR spectral decomposition tool \citep{Smith07a}. The flux of \htwo\ rotational lines for Haro\thinspace11 are from \citet{Cormier14}, while for UM\thinspace311, and CGCG\thinspace007-025 \citet{Hunt10} measured the \htwo\ rotational line fluxes. For Hubble V region, \citet{Roussel07} derived the flux of \htwo\ rotational lines. Figure \ref{fig:Haro} is an excitation diagram for low metallcity galaxy Haro\thinspace11, with the power law model fit. Our model prediction for the unobserved S(0) line is included, and is consistent with the estimated upper limit. The \htwo\ total gas mass for each low metallicity galaxy is measured using the power law model with lower temperature extrapolation to $T_{\ell}^{\star}$ = 49 K. Table \ref{table:lowz} lists the value of metallicity with their distance and the measured value of \htwo\ rotational line fluxes with the power law index and the \htwo\ gas mass derived using CO, dust, and our model. \begin{figure} \includegraphics*[width=0.5\textwidth]{fig13.eps} \caption{Excitation diagram for low metallicity galaxy Haro\thinspace11. The N$_{u}$/g$_{u}$ ratios are normalized with respect to S(1) transition. The dashed red line indicate the model fit to the observed ratios. The blue dashed line predicts the value of S(0) flux ratio. The model estimated $T_{\ell}$, and n are mentioned.} \label{fig:Haro} \end{figure} Figure~\ref{fig:lowZ} compiles \htwo\ gas masses derived using the various indirect tracers, together with the results from our \htwo-only model. All values are shown \emph{relative} to the \htwo\ mass inferred using CO luminosities with $\rm{\alpha_{CO} = \alpha_{CO,Gal}}$, and are plotted as a function of metallicity. For the SINGS sample, a variation of about 2--3 times the $\rm{L_{CO}}$-derived \htwo\ gas mass is found, likely a consequence of the intrinsic variation in $\rm{\alpha_{CO}}$ at high metallicity \twelveoh\ $\gtrsim 8.4$. At intermediate metallicity, the molecular gas content from CO line emission for the Hubble V region in NGC\thinspace6822 compares well with our model-derived gas mass, which suggests a similar Galactic conversion factor ($\rm{\alpha_{CO}}$ = $\rm{\alpha_{CO,Gal}}$), consistent with the results of \citet{Rijcke06}. At lower metallicity, however, our model disagrees strongly with naive CO-based estimates, yielding up to $\sim100\times$ the molecular gas mass inferred from CO emission. The molecular masses we recover at low metallicity are in good agreement with other measures which attempt to account for the impact of reduced metal abundance. These include dust-derived measurements, where available (when the strong metallicity dependence of the dust-to-gass ratio is accounted for). They also agree well with the prescription for the power-law like $\rm{\alpha_{CO}}$ variations with metallicity recovered from inverting star formation densities among all non-starburst galaxies in the HERACLES sample \citep{Schruba12}. The theoretical model of varying $\rm{\alpha_{CO}}$ by \citet{Wolfire10}, assuming \htwo\ column density of 10$^{22}$ cm$^{-2}$ in a molecular cloud, agrees well with observational results based on dust-mass at low metallicity \citep{Leroy11,Sandstrom13}. Adopting a solar metallicity value of 8.66 and heating rate/H atom $\log(G_0/n)$ = -0.3 (from the $\rm{L_{OI}/L_{FIR}}$ studies of \cite{Malhotra01}), the \citet{Wolfire10} model also agrees well with our \htwo\ rotational emission modeled masses. The power law model reliably recovers the total molecular gas masses at metallicity as low as 10$\%$ of the Milky Way, where CO and other indirect tracers suffer strong and non linear biases. \begin{figure*} \centering \includegraphics[width=0.95\textwidth]{fig14.eps} \caption{The ratio of molecular gas masses estimated using different methods to the ``naive mass'' obtained using $\rm{L_{CO}}$ (and $\rm{\alpha_{CO}}$ = $\rm{\alpha_{CO,Gal}}$), as a function of gas-phase metallicity. This ratio is equivalent to $\alpha_\mathrm{CO}/\rm{\alpha_{CO,Gal}}$. The black points show the molecular gas masses derived from the \htwo\ rotational model. The blue line shows a fit to this ratio derived from inverting the star formation law for HERACLES non-starburst galaxies \citep{Schruba12}. The molecular gas masses traced by the dust emission are denoted by the red points in the plot, and are shifted slightly in their metallicity values for clarity. The black solid line is the predicted mass ratio from the theoretical model of \citet{Wolfire10}, assuming $N({\textrm{H}_2})=10^{22}$ cm$^{-2}$ and $log(G_0/n)$ = -0.3. A steep increase in the ratio of model derived \htwo\ gas mass to the $\rm{L_{CO}}$ derived measurements is observed at metallicity values \twelveoh $\leq$ 8.4.} \label{fig:lowZ} \end{figure*} \subsection{Future prospects} At metallicities $\lesssim$ 0.25 Z$_{\odot}$, the direct power-law method recovers total molecular gas content as reliably as other tracers that account for or avoid the impact of reduced gas-phase metal content. At even lower metallicities, the CO abundance plummets, with essentially all of the molecular gas in a CO-dark phase. In the first epoch of the star formation in the universe, the extremely low abundance of heavy elements leaves \htwo\ as a principal coolant \citep{Lepp02}. The Mid Infrared Instrument (MIRI) onboard \emph{JWST} is sensitive enough to detect the S(1), S(2), S(3) and higher rotational lines of \htwo\ in luminous galaxies (U/LIRGs) till redshift 0.6, 1.3, 1.9, and higher, respectively. Assuming average $\rm{L_{H_{2}S(1)}/L_{IR}}$ = 10$^{-4}$ \citep{Bonato15}, and $\rm{L_{IR}}$ = 3$\times$10$^{11}$ L$_{\odot}$ (typical for LIRGs), for S(1) at z = 0.5 to have signal to noise S/N = 5 will require 30 minutes of integration time with JWST-MIRI. It will be possible to measure pure \htwo\ rotational lines at high redshifts of z $\approx$ 6--7, almost reaching the reionization era of universe, with the SPace Infrared telescope for Cosmology and Astrophysics (SPICA) and the Cryogenic Aperture Large Infrared Submillimeter Telescope Observatory (CALISTO), planned for the 2020 decade \citep{Roelfsema12, Bradford15}. The above mentioned future projects for the next decade provides an opportunity to observe \htwo\ rotational lines at high redshifts. The power law model can be an useful tool in estimating molecular gas mass and study its variation and consequences at different redshifts.

1607.08036

1607

1607.02322_arXiv.txt

{Among the candidates for generating turbulence in accretion discs in situations with low intrinsic ionization the vertical shear instability (VSI) has become an interesting candidate, as it relies purely on a vertical gradient in the angular velocity. Existing numerical simulations have shown that $\alpha$-values a few times $10^{-4}$ can be generated.} {The particle growth in the early planet formation phase is determined by the dynamics of embedded dust particles. Here, we address in particular the efficiency of VSI-turbulence in concentrating particles in order to generate overdensities and low collision velocities.} {We perform three-dimensional (3D) numerical hydrodynamical simulations of accretion discs around young stars that include radiative transport and irradiation from the central star. The motion of embedded particles within a size range of a fraction of mm up to several m is followed using standard drag formula.} {We confirm that under realistic conditions the VSI is able to generate turbulence in full 3D protoplanetary discs. The irradiated disc shows turbulence within 10 to 60 \AU. The mean radial motion of the gas is such that it is directed inward near the midplane and outward in the surface layers. We find that large particles drift inward with the expected speed, while small particles can experience phases of outward drift. Additionally, the particles show bunching behaviour with overdensities reaching 5 times the average value, which is strongest for dimensionless stopping times around unity. } {Particles in a VSI-turbulent discs are concentrated in large scale turbulent eddies and show low relative speeds that allow for growing collisions. The reached overdensities will also allow for the onset streaming instabilities further enhancing particle growth. The outward drift for small particles at higher disk elevations allows for the transport of processed high temperature material in the Solar System to larger distances.}

To drive mass flow in accretion discs an anomalous source of angular momentum is required \citep{2002apa..book.....F}. A strong candidate is the magneto-rotational instability (MRI), which gives rise to turbulent magnetohydrodynamical (MHD) flows that create an outward angular momentum transport discs \citep{1998RvMP...70....1B}. Driven by magnetic fields, the MRI requires a sufficient level of ionization to sustain a turbulent state within the disc. However, protoplanetary discs have only a very low temperature regime and insufficient thermal ionization. Even considering external sources of ionization there appears to be a region of insufficient ionization level such that the MRI cannot operate, as shown by resistive MHD simulations including radiative transport \citep{2012MNRAS.420.2419F}. Hence, there may exist a dead zone somewhere between $2-20\AU$ \citep{Armitage2011ARA&A..49..195A}, where the MRI can only produce very weak turbulence. Recent simulations which included, in addition to Ohmic resistivity, also ambipolar diffusion have even shown no signs of turbulence at all in this region \citep{2015ApJ...801...84G}. The Hall effect creates strong winds in the surface of the disc and may even reintroduce angular momentum transport in the dead zone, but this depends on the sign of the magnetic field \citep{Bai2014ApJ...791..137B,Bai2015ApJ...798...84B}. Thus another origin of instability inside the dead zones is warranted to drive accretion in protoplanetary discs. As an alternative to the MRI, different examples of purely hydrodynamic instabilities in discs have been suggested, such as the gravitational instability \citep{1987MNRAS.225..607L}, the convective instability \citep{1988ApJ...329..739R}, or the baroclinic instability \citep{2003ApJ...582..869K}, but they do not operate under general conditions. One possibility, that has attracted recently more attention is the vertical shear instability (VSI) suggested for accretion discs by \citet{Urpin2003A&A...404..397U}. The mechanism was first examined in relation to differential rotating stars \citep{Goldreich1967,1968ZA.....68..317F}, and it is also known as Goldreich-Schubert-Fricke instability. In the context of discs first simulations have been carried out by \citet{2004A&A...426..755A}. While there is a much larger radial gradient in the angular velocity, $\Omega$, to feed instabilities, most instabilities cannot overcome the stabilising effect of rotation. In the context of the VSI, it is the vertical shear in $\Omega$ created by a radial temperature gradient that allows the disc to become unstable. The numerical work of \citet{Nelson2013MNRAS.435.2610N} showed that a small turbulent $\alpha$-value in the range of a few $10^{-4}$ was possible for isothermal discs. For non-isothermal discs \citet{Nelson2013MNRAS.435.2610N} point out that radiative cooling (diffusion) and viscosity will reduce the instability, and they developed a theoretical model describing the initial vertical elongated modes destabilising the disc. In simulations that included full radiative transport \citet{Stoll2014A&A...572A..77S} showed that for situations typical in protoplanetary discs a sustained VSI was possible providing an $\alpha \sim 10^{-4}$. They found the development of a global wave pattern within the disc whose wavelength was determined partly by viscous effects. Later, \citet{Barker2015MNRAS.450...21B} analyzed the VSI through linear analyses of locally isothermal discs and support the modal behaviour seen in the non-linear simulations by \citet{Nelson2013MNRAS.435.2610N} and \citet{Stoll2014A&A...572A..77S}. They also stress the importance of viscosity to set the smallest length scale. Recently, \citet{Lin2015ApJ...811...17L} shed light on the cooling requirements of the VSI, that arise because the VSI has to compete with the stabilising vertical buoyancy. Their theoretical models predict activity in regions with large cooling time only for very large wavenumbers. Thus the VSI is limited by the cooling time on large scales and by viscosity on small scales. With this in mind they predict VSI activity for typical disc models only between 5 and 100 au. We test this idea by expanding our previous work \citep{Stoll2014A&A...572A..77S} where we used a self-consistent radiation transport module and a vertically irradiated disc in the simulations. Here, we treat the irradiation in a more realistic way as originating from the central star, and show that even under this condition the VSI can be sustained. The turbulence in protoplanetary discs is also critical for the initial dust evolution that leads eventually to planet formation. Numerical simulations performed by \citet{Johansen2005a} and \citet{Fromang2005MNRAS.364L..81F} showed that particles can be caught in the local pressure maxima generated by the MRI turbulence. This clustering of dust particles can then trigger a streaming instability \citep{2005ApJ...620..459Y} that will lead to further clustering and subsequently to the formation of $km$-planetesimals. In the context of the VSI the large scale velocity patterns of the corrugation mode promises interesting behaviour for embedded dust grains and larger particles. To investigate the impact of the VSI modes on the dust particles we add particles into our disc model and follow their dynamical evolution. The paper is organized as follows. In Sect.~\ref{sect:setup} we present our numerical and physical setup. We present a detailed analysis of an isothermal disc model in Sect.~\ref{sect:isodisc}, and we discuss in Sect.~\ref{isothermal} the results for the particle evolution is this model. In Sect.~\ref{sec:viscosity} we describe the results of a viscous model. The simulations with radiation transport and stellar irradiation are presented in Sect.~\ref{radiative} and in Sect.~\ref{conclude} we conclude.

In the paper we analyzed the dynamics of particles embedded in hydrodynamic discs that show fully developed turbulence as induced by the VSI. In a first step we calculated isothermal disc models in full three dimensions and analyzed the properties of the turbulence generated by the VSI. Our standard model consisted of an eighth of a full circle ($\phi_{max} = \pi/4$) and showed in the fully developed turbulent state $\alpha$-values around $6 \cdot 10^{-4}$, which is of the same order of magnitude or even slightly larger than the corresponding 2D models \citep{Stoll2014A&A...572A..77S}. The 3D models shows variations in the azimuthal direction and these fluctuations follow a Kolmogorov-type spectrum. The mean radial velocity of the gas in a VSI turbulent disc turned out to be directed inward in the disc midplane and outward in the upper layers, in agreement with global MHD simulations using zero net vertical magnetic flux \citep{2011ApJ...735..122F}. This flow is opposite to viscous laminar discs \citep{1984SvA....28...50U,1992ApJ...397..600K} or MHD discs with non-zero vertical magenetic field \citep{2014ApJ...784..121S}. For 3D discs covering the full circle ($\phi_{max} = 2 \pi$) we found very similar results, which allowed us to treat particle evolution in the reduced domain. In addition to the isothermal case we studied fully radiative models including heating from the central star. To allow for regimes where the VSI instability can operate we extended to radial domain from $8-80$\AU. The temperature structure in the disc displayed a central disc region with a nearly constant temperature in the vertical direction and hotter surface layers produced by the stellar irradiation. The vertically varying opacity in the disc resulted in different cooling times and the turbulence turned out to be slightly weaker in comparison to the purely isothermal situation. For the effective $\alpha$-parameter values of around $10^{-4}$ were reached in the active state that extended from about 10 to 60 \AU. After having reached the equilibrium state we inserted particles of different sizes to study their motion in the disc, where the drag force between gaseous disc and particles was treated in the Epstein regime. Overall we found for both, isothermal and radiative discs comparable results. On average the particles drift inwards with the expected speed. For all disc models we found that the smallest particles show an outwardly directed radial drift. This comes about because the small particles are coupled more to the gas flow and are lifted upward by the vertical motions of the VSI induced large scale flows. Since the average flow direction in the upper layers is positive small dust particles that are elevated above the disk's midplane are dragged along and move outwards. Particles below about 1~mm in size experience this fate. This outward drift might be beneficial in transporting strongly heated solid material to larger radii as required to explain for example the presence of chondrules at larger radii in the Solar System \citep{2002A&A...384.1107B}. The upward drift of small particles in the disc by the VSI modes will also help to explain the observed presence of a population small particles in the later stages of the disc evolution that were produced by a fragmentation process \citep{2005A&A...434..971D}. Using the information of histograms, probability functions and pair correlation functions we analyzed the spatial re-distribution of particles in the disc that were initially homogeneously distributed. We found that the particles are strongly 'bunched' together by the large scale motions of the VSI turbulence. The bunching effect is strongest for particles with a stopping time of the order unity and the maximum overdensities reached were about 5 times the average initial density of the particles. The relative velocity between particles of the same size is smallest (about a few m/s) for those particles that show the strongest bunching. This combination of high density and low relative speed is highly beneficial for the early formation process of planetary precursors. First, at these relative speeds collisions between two particles can lead to sticking collisions \citep{2008ARA&A..46...21B,2013MNRAS.435.2371M}. The higher relative velocities between particles of different sizes does not necessarily lead to fragmentation. The experiments of \citet{2009MNRAS.393.1584T} have shown that particles with different size can stick to each other even for collisions up to 50 m/s and possibly more. Secondly, through the concentration of particles it is possible to trigger streaming instabilities in the disc which can further increase the particle concentration and growth \citep{2005ApJ...620..459Y}. The two dimensional distribution of particles in the disc shows axisymmetric ring-like concentration zones of the particles resembling very roughly the features observed recently in the disc around HL~Tau \citep{Brogan2015ApJ...808L...3A}. Even though the strongest effect is seen here in our simulations for particles about one meter in size, it is possible that through collisions of nearly equal sized bodies much smaller particles that could generate the observed emission can be produced and which follow a similar spatial distribution. Obviously the observed spacing of the 'bright' rings in our simulations is smaller than those observed in HL~Tau but the inclusion of variations in opacity or chemical abundances may create larger coherent structures.

1607.02322

1607

1607.05734_arXiv.txt

{The consistency of planet formation models suffers from the disconnection between the regime of small and large bodies. This is primarily caused by so-called growth barriers: the direct growth of larger bodies is halted at centimetre-sized objects and particular conditions are required for the formation of larger, gravitationally bound planetesimals.} {We aim to connect models of dust evolution and planetesimal formation to identify regions of protoplanetary discs that are favourable for the formation of kilometre-sized bodies and the first planetary embryos.} {We combine semi-analytical models of viscous protoplanetary disc evolution, dust growth and drift including backreaction of the dust particles on the gas, and planetesimal formation via the streaming instability into one numerical code. We investigate how planetesimal formation is affected by the mass of the protoplanetary disc, its initial dust content, and the stickiness of dust aggregates.} {We find that the dust growth and drift leads to a global redistribution of solids. The pile-up of pebbles in the inner disc provides local conditions where the streaming instability is effective. Planetesimals form in an annulus with its inner edge lying between 0.3~AU and 1~AU and its width ranging from 0.3~AU to 3~AU. The resulting surface density of planetesimals follows a radial profile that is much steeper than the initial disc profile. These results support formation of terrestrial planets in the solar system from a narrow annulus of planetesimals, which reproduces their peculiar mass ratios.} {}

Planets seem to be omnipresent in our Galaxy with most stars orbited by one or more of them \citep{2012Natur.481..167C}. However, we are a long way from a complete understanding of how these planets form. Despite recent progress in planet formation research, there is no model that can explain the growth of planets beginning with micron-sized dust grains. Most significantly, there is a disconnection between the models of the early stages of planet formation that are dealing with dust growth, and the late stages which follow the final accretion of planetary systems starting with planetesimals and embryos. The late stage models usually assume that planetesimals form rapidly with a smooth radial distribution \citep{2013apf..book.....A}. On the other hand, the early stage models typically end without producing any aggregates larger than cm-sized \citep[see e.g.][]{2014prpl.conf..339T}. This is because of so-called growth barriers that result from the collisional physics of dust aggregates and loss of solids because of their radial drift. The dust aggregates tend to bounce and fragment at the impact speeds predicted for the protoplanetary disc environment \citep{2010A&A...513A..57Z,2011A&A...525A..11B}. On the other hand, the radial drift timescale is shorter than the growth timescale for aggregates approaching centimetre sizes and the initial aggregates are removed from the disc before any larger bodies can form \citep{2008A&A...480..859B,2009A&A...503L...5B}. We give a concise overview of the typical predictions for dust evolution in the following section. This paper is organised as follows. We give a brief introduction to the problem of growth barriers in Sect.~\ref{sub:barriers} and highlight the streaming instability as a possible solution in Sect.~\ref{sub:si}. In Sect.~\ref{sub:methods}, we explain the numerical approach that we employ to investigate dust evolution and planetesimal formation in a viscously evolving protoplanetary disc. We present results of our models in Sect.~\ref{sub:results}. Finally, we discuss limitations of our approach in Sect.~\ref{sub:discussion} and conclude the work in Sect.~\ref{sub:last}. For the readers convenience, the symbols used throughout this paper are summarised in Table~\ref{table:symbols}. \subsection{Growth barriers}\label{sub:barriers} \begin{figure*} \centering \includegraphics[width=0.8\hsize]{map.png} \caption{Overview of the growth barriers in our fiducial disc model. The red line corresponds to a Stokes number of unity. The yellow region marks the fragmentation barrier where the impact speeds would be higher then the threshold value of 10~m~s$^{-1}$. The grey region shows the size scale for which the drift timescale is shorter than the growth timescale, meaning that the aggregates would be removed from a given location faster than the growth could replenish them: this is the drift barrier. The light grey vertical lines are tracks of test particles evolving by growth and drift in a steady-state disc. Finally, the blue nearly horizontal line shows the size scale corresponding to the Stokes number of 10$^{-2}$, which is a minimum size required for planetesimal formation via the streaming instability.} \label{fig:barriersmap} \end{figure*} Dust evolution is driven by its interaction with the sub-Keplerian gas disc. The interaction between a dust aggregate and the surrounding gas can be conveniently parametrised with the so-called Stokes number \begin{equation}\label{stokes1} {\rm{St}} = t_{\rm{s}}\Omega_{\rm{K}}, \end{equation} where $t_{\rm{s}}$ is the timescale over which the aggregate adjusts its velocity to the velocity of surrounding gas, and $\Omega_{\rm{K}}$ is Keplerian frequency. In this way, the Stokes number is the ratio between how quickly the aggregate reacts to the gas and the orbital period. Particles with $\rm{St}\ll1$ are tightly coupled to gas and $\rm{St}\gg1$ indicates decoupled solids. For our needs, it is convenient to rewrite Eq.~(\ref{stokes1}) as \begin{equation}\label{stokes} {\rm{St}} = \frac{\pi}{2}\frac{a\rho_{\bullet}}{\Sigma_{\rm{g}}}, \end{equation} where $a$ is the radius and $\rho_{\bullet}$ is the internal density of the dust aggregate, and $\Sigma_{\rm{g}}$ is the surface density of the gas. Equation~(\ref{stokes}) applies to small and compact solids, where the radius $a$ is smaller than the mean-free path in gas. This is the so-called Epstein drag regime, where particles are well-coupled to the gas, and this regime is valid for all models presented in this paper. Figure~\ref{fig:barriersmap} gives an overview of global dust evolution in a protoplanetary disc with total mass of $0.1$~M$_{\odot}$, in terms of distance to the central star and dust aggregate size. The size of aggregates corresponding to the Stokes number of unity, which is when they are affected by the interaction with the gas the most, is marked with the red nearly horizontal line. The blue line below shows a minimum size that is required for an efficient streaming instability (corresponding to ${\rm{St}}=10^{-2}$, see Sect.~\ref{sub:si}). The outer part of the protoplanetary disc is dominated by the radial drift barrier (grey triangle-shaped region), where the growth timescale \begin{equation}\label{tgrowth} \tau_{\rm growth} = a~\left(\frac{{\rm d}a}{{\rm d}t}\right)^{-1} = \frac{a\rho_{\bullet}}{\rho_{\rm{d}} \Delta v} \end{equation} is longer than the radial drift timescale \begin{equation}\label{tdrift} \tau_{\rm drift} = r \left(\frac{{\rm d}r}{{\rm d}t}\right)^{-1} = \frac{r}{|v_{\rm r,d}|}, \end{equation} where $\rho_{\rm{d}}$ is the density of dust, $\Delta v$ is the impact velocity for collisions between dust aggregates, $r$ is the radial distance to the star, and $v_{\rm r,d}$ is the radial drift velocity. We assume that the growth preferentially happens between equal-sized grains that have settled to the midplane when deriving Eq.~(\ref{tgrowth}). This means that the dust density may be written as \begin{equation}\label{rhod} \rho_{\rm{d}} = \frac{\Sigma_{\rm{d}}}{\sqrt{2\pi}H_{\rm{d}}}, \end{equation} where $\Sigma_{\rm{d}}$ is the dust surface density and $H_{\rm{d}}$ is the scale height of the dust, which depends on the turbulence strength $\alpha_{\rm{t}}$ and the Stokes number of grains $\rm{St}$. In the case of a single-sized population \begin{equation}\label{Hd} H_{\rm{d}} = H_{\rm{g}} \sqrt{\frac{\alpha_{\rm{t}}}{\alpha_{\rm{t}}+{\rm{St}}}}, \end{equation} where $H_{\rm{g}} = c_{\rm{s}} \Omega_{\rm{K}}^{-1}$ is the scale height of the gas. We note that the growth timescale $\tau_{\rm{growth}}$ given by Eq.~(\ref{tgrowth}) simplifies to \begin{equation}\label{tgrowth2} \tau_{\rm{growth}} \approx \frac{1}{Z\cdot\Omega_{\rm K}} \end{equation} under the assumptions that the collisions are driven by turbulence, in which case the impact speed $\Delta v \approx \sqrt{3\alpha_{\rm t}{\rm St}}\cdot c_{\rm s}$ \citep{2007A&A...466..413O}, the dust grains are in the Epstein regime (see Eq.~\ref{stokes}), and the dust density in the midplane is given by Eq.~(\ref{rhod}). We emphasise that, because $\Delta v \propto {\alpha_{\rm t}}^{1 / 2}$ and $\rho_{\rm d} \propto {\alpha_{\rm t}}^{-1 / 2}$, the dependence of the dust growth timescale on the turbulence strength parameter $\alpha_{\rm t}$ cancels itself out. This initial growth stage is significant particularly at large orbital distances, where the growth is much slower than in the inner parts of the protoplanetary disc. This results in a delayed delivery of pebbles from the outer disc to its inner regions. The same effect was also described by \citet{2014A&A...572A.107L}, who called it a "pebble formation front". The radial drift limits the maximum size of grains that are available in the outer parts of the disc. The pebbles that grow far away from the star are then shifted to the inner disc, where they undergo fragmentation during high-speed collisions. We assume that the fragmentation happens when the collision speed exceeds a threshold value, which we set to $v_{\rm{th}}=10$~m~s$^{-1}$ for Fig.~\ref{fig:barriersmap}. The evolution of dust in the outer disc is determined by the interplay between growth and drift. Because the timescale of drift is shorter than what is needed to grow to centimetre sizes at a few tens of AU from the central star, this outer region is gradually depleted on a timescale of a few Myrs. However, as may be seen in Fig.~\ref{fig:barriersmap}, the radial drift barrier does not stretch all the way to the inner edge of the disc. In the inner disc, the collisional timescale is shorter than the drift timescale. If growth happens even at high impact speeds, for example in the case of very porous dust aggregates, planetesimal formation via direct growth may be possible \citep{2012ApJ...752..106O, 2013A&A...557L...4K}. Even for compact grains, interplay between the radial drift and dust growth can lead to a pile-up of solids in the inner regions of the disc \citep{2012A&A...539A.148B, 2012A&A...537A..61L, 2014A&A...565A.129P}. This pile-up is required by the planetesimal formation models which include the streaming instability \citep{2009ApJ...704L..75J,2010ApJ...722.1437B}. At the same time, the streaming instability requires the presence of relatively large pebbles, with sizes corresponding to the Stokes number of $\rm{St}>10^{-2}$ \citep{2014A&A...572A..78D}. This is not necessarily the case because of the bouncing and fragmentation barriers. However, as can be seen in Fig.~\ref{fig:barriersmap}, there is a region around 1-10~AU where the maximum size of grains that can be reached because of fragmentation is above the $\rm{St}=10^{-2}$ line and the radial drift barrier is not efficient. In this paper, we check whether the redistribution of solids driven by the radial drift and growth in a realistic viscous disc can lead to planetesimal formation via the streaming instability. \subsection{Streaming instability}\label{sub:si} Owing to the growth barriers described in the previous section, direct growth from micron to km-sizes seems unlikely. Streaming instability that is able to produce planetesimals directly out of cm-sized pebbles is a good solution to the planetesimal formation issue \citep{2007Natur.448.1022J}. However, planetesimal formation via the streaming instability requires enhancement of the dust-to-gas ratio by a factor of a~few over the standard solar value of 10$^{-2}$ \citep{2009ApJ...704L..75J, 2010ApJ...722.1437B}. There are different scenarios that modify disc structure and introduce pressure bumps that make it possible to obtain such enhancements: the zonal flows \citep{2011A&A...529A..62J, 2013ApJ...763..117D}, vortices \citep{2015ApJ...804...35R, 2016arXiv160105945S}, dead zone edges \citep{2009A&A...497..869L}, and planet-disc interactions \citep{2012MNRAS.423.1450A}. In the models presented in this paper, we check whether the streaming instability can work thanks to a dust pile-up induced in the inner disc by the combination of growth and drift. This scenario does not require any ad-hoc assumptions about the disc structure or pre-existing planets. Until now the streaming instability was only modelled in local or quasi-global simulations. We use a global 1D semi-analytical protoplanetary disc model together with a prescription for streaming instability extracted from the local hydrodynamic simulations, similar to \citet{2014A&A...572A..78D} for a local case. As a consequence, we are able to identify regions of the disc in which planetesimal formation happens. We perform an extended parameter study to investigate how the planetesimal formation is affected by the disc mass, metallicity, and stickiness of dust aggregates. The modelling methods that we use are described in Sect.~\ref{sub:methods}.

\label{sub:last} This paper addresses the connection between dust evolution and planetesimal formation, which represents a major gap in state-of-the-art planet formation models. As dust growth is limited by fragmentation and radial drift, a direct growth to planetesimal sizes seems unlikely. However, the same processes drive a global redistribution of solids and may lead to a pile-up of pebbles that triggers planetesimal formation via the streaming instability. We show that a narrow ring of planetesimals with a steep surface density profile is naturally produced in the inner part of a shallow protoplanetary disc corresponding to the observed ones. These planetesimals form from the flux of solids originating from the outer disc that are carried by the radial drift. At the same time, the radial drift limits the radial extent of this annulus, as it prevents the dust-to-gas ratio enhancement from spreading further out. On the other hand, the inner edge is caused by lack of sufficiently large pebbles that could trigger the streaming instability. Impact velocities increase towards the inner disc because of higher temperatures and thus the maximum size of grains decreases. The exact properties of planetesimal annulus depend on various parameters, as shown in Fig.~\ref{fig:4panels} and summarised in Table~\ref{table:allruns}. The narrow annulus of planetesimals around 1~AU enables us to reproduce the unusual masses of terrestrial planets, where Earth and Venus are significantly more massive than Mercury and Mars \citep{2009ApJ...703.1131H}. In some of our models, the planetesimal annulus contains as much as 1 000 Earth masses, which could allow for in situ formation of the close-in massive planets that have been detected around many stars \citep{2013ApJ...766...81F}. Our results, including the surface density of planetesimals and timing of their formation, may be used as an input to models which investigate the later stages of planet accretion, planet population synthesis, and the internal evolution of asteroids. The size and flux of pebbles that we obtain is important for the pebble accretion models that typically assume unphysically large aggregates and make ad hoc assumptions about their surface density.

1607.05734

1607

1607.05996_arXiv.txt

During the common envelope (CE) phase, a giant star in a binary system overflows its Roche lobe and unstable mass transfer leads to a spiral-in of the companion, resulting in a close binary system or in a merger of the stellar cores. Dynamo processes during the CE phase have been proposed as a mechanism to generate magnetic fields that are important for forming magnetic white dwarfs (MWDs) and for shaping planetary nebulae. Here, we present the first magnetohydrodynamics simulations of the dynamical spiral-in during a CE phase. We find that magnetic fields are strongly amplified in the accretion stream around the $1M_\odot$ companion as it spirals into the envelope of a $2M_\odot$ RG\@. This leads to field strengths of 10 to 100\,kG throughout the envelope after 120\,d. The magnetic field amplification is consistent with being driven by the magnetorotational instability. The field strengths reached in our simulation make the magnetic field interesting for diagnostic purposes, but they are dynamically irrelevant. They are also too small to explain the formation of the highest fields found in MWDs, but may be relevant for luminous red novae, and detecting magnetic fields in these events would support the scenario as proposed here.

\label{sec:introduction} The CE phase is usually invoked to explain the formation of close binary systems with at least one compact star: the compact star is born as a core of a giant with a much larger radius than the separation of the observed binary. During the CE phase, angular momentum and energy is extracted from the system as the envelope is ejected and a close binary system emerges. Results of this evolution include cataclysmic variables \citep[CVs,][]{king1988a,ritter2003a}, close WD and main sequence (MS) binaries \citep{schreiber2003a,zorotovic2010a}, double WDs \citep{iben1985a,han1995a,nelemans2001a}, Type Ia supernovae \citep{iben1984a,ruiter2009a,toonen2012a}, and many more \citep[see also the reviews by][]{iben1993a,taam2000a,ivanova2013a}. The origin of magnetic fields in stars is still unknown, and it is debated if they are remainders from fields generated during the formation of a star (fossil field hypothesis) or if they are generated during the life of a star by dynamo processes \citep{ferrario2015a}. WDs show magnetic fields with strengths up to $10^9$\ G and can also occur in binaries with accretion from a low-mass companion, so-called magnetic CVs \citep{ferrario2015b}. The absence of magnetic WDs in wide binaries suggests a binary origin for both magnetic WDs and magnetic CVs \citep{tout2008b,briggs2015a}. In this scenario, magnetic fields are amplified by a dynamo process in the differentially rotating envelope during a CE phase \citep{regos1995a} or in an accretion disk around the giant core that was formed from the tidally disrupted companion \citep{nordhaus2011a}. The magnetic fields created during the CE phase have to be much larger than the final surface field of the WD because it is difficult to anchor the fields on the WD \citep{potter2010a}. At the high-mass end of magnetic WDs, mergers of two WDs constitute a different channel that can explain large magnetic fields \citep{zhu2015a}. In their population study, \citet{briggs2015a} find that the major contribution comes from mergers during the CE phase. Magnetic fields generated during the CE phase may also affect the shape and evolution of planetary nebulae (PNe) \citep{nordhaus2006a,nordhaus2007a}. The magnetic field generation relies again on a dynamo process in the differentially rotating envelope or in an accretion disk formed by the tidally disrupted companion. Magnetic fields have not been detected in post-AGB stars as remnants of PNe; the upper limits lie at 100 to 300 G \citep{jordan2012c}. \citet{tocknell2014a} used observational data on four jets in PNe to constrain the magnetic fields that are necessary to launch these jets via the Blandford--Payne mechanism \citep{blandford1982a} to hundreds of G to a few kG\@. Three jets pre-date the CE event by few thousand years, one jet post-dates the CE event by a few thousand years. Thus, besides binarity itself \citep{demarco2009a}, magnetic fields can be an important factor for shaping PNe. Up to now, the generation of magnetic fields during the CE phase was attributed to dynamo processes driven by shear due to differential rotation in the envelope or in an accretion disk. Here, we present the first magnetohydrodynamics simulations of the dynamical spiral-in during a CE phase. The simulations extend the hydrodynamics simulation of a CE phase of a $2M_\odot$ RG interacting with a $1M_\odot$ companion presented by \citet{ohlmann2016a} and account for MHD effects.

\label{sec:discussion} The first MHD simulations of the CE phase show that magnetic fields are strongly amplified in the direct vicinity of the companion star on dynamical timescales during the spiral-in. The temporal and spatial scales are compatible with the MRI operating in the accretion flow around the companion. Although the magnetic field strength is increased by orders of magnitude, the dynamical impact is small: mass loss is slightly increased during the first orbit (about 5\% to 6\%), but the contribution to angular momentum and energy transport is not significant and the final separation of the stellar cores is very similar to that found in non-MHD simulations with the same initial parameters. This is a new way of generating magnetic fields during the dynamical spiral-in of the CE phase compared to earlier investigations that assumed a dynamo operating in the differentially rotating envelope \citep{regos1995a}, in an accretion disk formed from the tidally disrupted companion \citep{nordhaus2011a}, or in the hot outer layers of the degenerate core \citep{wickramasinghe2014a}. These processes, however, may still be important later in the evolution. We also stress that we only simulated a single system, and that a variety of systems undergoing a CE phase has to be simulated to understand the range of magnetic fields that can be generated during CE evolution. Magnetic fields that vary on short timescales, as is the case in the simulations presented here, are difficult to anchor at the WD surface \citep{potter2010a}. Moreover, the magnetic fields present in the simulations (10 to 100\,kG) are smaller than the largest fields in WDs (up to $10^9$\,G, see \citealp{ferrario2015b}). Thus, the dynamical amplification of magnetic fields during the spiral-in may not explain the formation of high-field magnetic WDs for the simulated CE phase of a $2M_\odot$ RG\@. According to the population synthesis calculations by \citet{briggs2015a}, the largest contribution to high-field magnetic WDs comes from mergers during the CE phase of an AGB primary. Hence, such systems may be more promising for future studies. RGs can be observed with mean magnetic fields down to 10 to 100\,G, but all observed RGs with magnetic fields cluster at the base of the RG branch \citep{auriere2015a}. Because it is difficult to compute the photosphere in our simulations, the magnetic field at the photosphere cannot be predicted from our simulations, and we postpone such an analysis to future studies. Nevertheless, CE events have been connected to luminous red novae \citep[LRN,][]{ivanova2013b}, and the presence of magnetic fields in these events would support dynamical amplification during the spiral-in phase of a CE event. \vspace{-3mm}

1607.05996

1607

1607.07394_arXiv.txt

Lorentz symmetry violations can be parametrized by an effective field theory framework that contains both General Relativity and the Standard Model of particle physics, called the Standard-Model Extension or SME. We consider in this work only the pure gravitational sector of the minimal SME. We present new constraints on the SME coefficients obtained from lunar laser ranging, very long baseline interferometry, and planetary motions.

The solar system remains the most precise laboratory to test the theory of gravity, that is to say General Relativity (GR). Constraints on deviations from GR can only be obtained in an extended theoretical framework that parametrizes such deviations. The parametrized post-Newtonian formalism is one of them and has been widely used for decades. More recently, other phenomenological frameworks have been developed like the Standard-Model Extension (SME), which is an extensive formalism that allows a systematic description of Lorentz symmetry violations in all sectors of physics, including gravity. We present here new constraints on pure-gravity sector coefficients of the minimal SME obtained with very long baseline interferometry (VLBI), lunar laser ranging (LLR) and planetary motions. We also assess the possibility to constrain them using future asteroids observations by Gaia.

We presented our latest constraints on gravity-sector SME coefficients obtained with LLR and VLBI observations. We highlighted also the improvement that we can expect from Gaia observations of asteroids in the future. A combined analysis with planetary ephemerides analysis,\cite{2015PhRvD..92f4049H} Lunar Laser Ranging,\cite{2007PhRvL..99x1103B,2016arXiv160700294B} atom interferometry,\citep{2008PhRvL.100c1101M} and binary pulsars\cite{2014PhRvL.112k1103S} would also be very interesting in order to decorrelate almost all gravity-sector SME coefficients and produce the most stringent estimate on the SME coefficients. Our analysis needs to be extended to include gravity-matter Lorentz violation in the SME framework.

1607.07394

1607

1607.02787_arXiv.txt

{To investigate galaxy properties as a function of their total stellar mass, we obtained 21cm \HI\ line observations at the 100-m class \nan\ Radio Telescope of 2839 galaxies from the Sloan Digital Sky Survey (SDSS) in the Local Volume (900$<$$cz$$<$12,000 \kms), dubbed the \nan\ Interstellar Baryons Legacy Extragalactic Survey (NIBLES) sample. They were selected evenly over their entire range of absolute SDSS $z$-band magnitudes (\Mz\ $\sim$ $-$13.5 to $-$24 mag), which were used as a proxy for their stellar masses. Here, a first, global presentation of the observations and basic results is given; their further analysis will be presented in other papers in this series. The galaxies were originally selected based on their properties, as listed in SDSS DR5. Comparing this photometry to their total \HI\ masses, we noted that, for a few percent, the SDSS magnitudes appeared severely misunderestimated, as confirmed by our re-measurements for selected objects. Although using the later DR9 results eliminated this problem in most cases, 384 still required manual photometric source selection. Usable \HI\ spectra were obtained for 2600 of the galaxies, of which 1733 (67\%) were clearly detected and 174 (7\%) marginally. The spectra for 241 other observed galaxies could not be used for further analysis owing to problems with either the \HI\ or the SDSS data. We reached the target number of about 150 sources per half-magnitude bin over the \Mz\ range $-$16.5 to $-$23 mag. Down to $-$21 mag the overall detection rate is rather constant at the $\sim$75\% level but it starts to decline steadily towards the 30\% level at $-$23 mag. Making regression fits by comparing total \HI\ and stellar masses for our sample, including our conservatively estimated \HI\ upper limits for non-detections, we find the relationship log(\MHIMstar) = $-$0.59 log(\Mstar) + 5.05, which lies significantly below the relationship found in the \MHIMstar\ - \Mstar\ plane when only using \HI\ detections. }

% Understanding the gas cycle in galaxies -- how galaxies acquire, process, and expel their gas -- is the central goal of most studies of galaxy evolution. Our current understanding is that this cycle is a balance between the accretion of gas onto the galaxy, the efficiency of turning the accreted and ``recycled'' gas into stars, and ejecting gas through a coupling of the gas to the luminous and mechanical energy output of stars and active galactic nuclei \citep{bouche10, lilly13}. However, well-studied galaxies, such as our own Milky Way, point to a very different picture \citep{haywood13}. The Milky Way's star formation rate has been roughly constant over the last 9 Gyrs and likely did not drive significant outflows during that period \citep{snaith14, lehnert14}. This is over a time span during which the cosmological accretion of dark matter was thought to decline by an order-of-magnitude \citep{neistein08,dekel09, dekel13}. To accommodate the high accretion rates onto galaxies relative to their star formation rates, studies often focus on ways of having galaxies drive vigorous massive outflows many times their star formation rates \citep{mitra15}. While this is logical, perhaps it is also important to search for processes that slow down the accretion timescale and the growth of the gas content of galaxies. One plausible way, which is certainly not unique, is to consider the growing angular momentum of accreted gas with decreasing redshift, which has the natural effect of increasing the timescale over which gas is made available for star formation \citep[e.g.,][]{lehnert15}. The role of \HI\ in galaxy formation and evolution is not yet completely clear. The reservoir of \HI\ gas in galaxies must ultimately feed their star formation \citep{vollmer11}, after cooling and forming molecular clouds. These gaseous disks are very extended, typically beyond the optically bright region of the galaxy \citep{bigiel12}. Since rotation curves are approximately flat out across these outer extended \HI\ disks \citep{vanalbada85}, they dominate the specific angular momentum budget of the galaxy, i.e., the angular momentum per unit mass. This is an important clue to their formation and their longevity. However, to interpret this important clue requires us to have a complete census of the \HI\ content of galaxies. To interpret spatially resolved observations of \HI\ disks, we need to place them into the general context of galaxies. Moreover, although integrated detections of galaxies (at any wavelength) provide only limited constraints on models of galaxy evolution, general demographics of galaxies and gas-phase distributions as a function of mass, environment, and morphological type, are at the moment the only characteristics that models are able to reliably predict. This is simply due to our rudimentary understanding of the physics underpinning galaxy evolution \citep{silk12}. There are two basic approaches to large \HI\ surveys of galaxies: blind surveys where the sky is scanned to search for detections, and pointed surveys targeting a high number of individual galaxies. Both approaches have their strengths and weaknesses. Blind surveys are best for unbiased detection of \HI-bearing galaxies, even discovering galaxies not previously known \citep{giovanelli13}, and for determining the unbiased comoving density of \HI\ in the local universe \citep[e.g.,][]{zwaan05, martin10}. The disadvantages are that most of the sky is free of \HI\ emission from galaxies, making the surveys time consuming and enabling them to reach only modest depths, and that a sample of \HI-selected galaxies will under-represent populations of galaxies that have low \HI\ content. Moreover, determining detection limits can be tricky and, almost by definition, the upper limits for undetected galaxies lie at similar \HI\ masses as the detections \citep[e.g.,][]{papastergis12}. Thus the upper limits do not add significantly to the analysis of global properties dependent on \HI\ mass, which negates some of the advantages of blind surveys. Pointed surveys have the advantage that the observed sample can be well-selected on particular properties, such as stellar mass or environment, are relatively economical since each pointing guarantees information, whether a detection or an upper limit, and are important for providing multi-variant information. The disadvantages of course are that pointed surveys can be biased in the galaxies they observe, leaving little room for important serendipitous discoveries. To aid in the determination of the \HI\ content of galaxies over a wide range of stellar masses, and overcome some of our remaining ignorance of the ``how much'' and ``where is'' of atomic gas, we undertook an \HI\ survey dubbed NIBLES, for \nan\ Interstellar Baryon Legacy Extragalactic Survey \citep{nibles08a, nibles08b, nibles09}. We observed 2850 galaxies in the local Universe (900$<$$cz$$<$12,000 \kms), selected as uniformly as possible on total stellar mass (for which we used the absolute $z$-band magnitude as a proxy) from the Sloan Digital Sky Survey (SDSS; see, e.g., \citealt{york00}). The data were obtained with the 100m-class \nan\ Radio Telescope (NRT; see Sect.~\ref{observations}). We subsequently supplemented it by four times more sensitive observations of over 150 objects at the 305m Arecibo radio telescope (see Sect.~\ref{observations}). NIBLES, with its uniform selection of galaxies based on total stellar mass, is aimed to complement other recent and/or ongoing large \HI\ surveys in the local volume. These surveys are, in order of the time at which they were started: \begin{enumerate} \item {HIPASS:} blind survey at the Parkes 64 m telescope \citep{barnes01}. Beam FWHM 14$'$, $rms$ noise level $13~\mathrm{mJy\,beam}^{-1}$ at a velocity resolution of 18 \kms, $-$90$^{\circ}$$<$$\delta$$<$25$^{\circ}$, search range $-$1280 to 12,700 \kms, data taken in 1997-2002 \citep{barnes01}. A total of $\sim$5300 galaxies were detected. The major galaxy catalogs are \citet{meyer04, wong06}; \item {ALFALFA:} blind survey at the Arecibo 305 m telescope. HPBW 4$'$, $rms$ $2.4~\mathrm{mJy\,beam}^{-1}$ at a velocity resolution of 10 \kms, 0$^{\circ}$$<$$\delta$$<$36$^{\circ}$, search range $-$2000 to 18,000 \kms, data taken in 2005-2012, not counting single-horn receiver follow-up observations \citep{giovanelli05a}. A total of $\sim$30,000 galaxies are expected to be detected. The first galaxy catalogs are \citet{giovanelli07, saintonge08, kent08, martin09, stierwalt09}, the subsequent $\alpha$.40 catalog \citep{haynes11} contains 15,855 detections over 40\% of the final survey area; the recently uploaded online $\alpha$.70 catalog (http://egg.astro.cornell.edu/alfalfa/data/) contains 25,534 detections over 70\% of the final survey area; \item {AGES:} blind survey at the Arecibo 305 m telescope of selected small ($\sim$$5^{\circ}\times$$5^{\circ}$) areas sampling different kinds of galaxy environments. HPBW \am{3}{5}, $rms$ $0.6~\mathrm{mJy\,beam}^{-1}$ at a velocity resolution of 10 \kms, search range $-$2000 to 20,000 \kms, data taking started in 2005. A total of 927 objects were detected so far. The galaxy catalogs are \citet{auld06, cortese08, irwin09, minchin10, davies11, taylor12, taylor13, taylor14a, taylor14b, minchin16, keenan16}; \item {GASS:} pointed survey at the Arecibo 305 m telescope \citep{catinella12} of 666 galaxies with stellar masses greater than 10$^{10}$ \Msun\ selected from the SDSS spectroscopic and the Galaxy Evolution Explorer (GALEX) ultraviolet imaging surveys. HPBW \am{3}{3}, mean $rms$ $0.74~\mathrm{mJy\,beam}^{-1}$ at a velocity resolution of 10-21 \kms, 6750$<$\Vopt$<$15,000 \kms, data taken in 2008-2012. A total of 379 galaxies were detected. The final galaxy catalogs are \citet{catinella10, catinella12, catinella13}. Note: not to be confused with the GASS survey at Parkes of Galactic \HI\ \citep{mcclure09}; \item {EBHIS:} blind survey at the Effelsberg 100 m telescope, of both Galactic and extragalactic sources \citep{winkel10,kerp11,winkel15}. HPBW \am{10}{8}, $-$5$^{\circ}$$<$$\delta$$<$90$^{\circ}$, search range $-$2000 to 18,000 \kms, data taking started in 2009. The current $rms$ for the extragalactic data is $23~\mathrm{mJy\,beam}^{-1}$ at a velocity resolution of 10 \kms\ \citep{floer14}, but observations for a second coverage of the Northern sky are underway which will lower the $rms$ to the level of HIPASS, so a similar sky density of \HI\ detections can be expected. \end{enumerate} We omitted the HIJASS blind survey at the 76 m Lovell Telescope at Jodrell Bank which yielded interesting early results, with 424 detections, but was never finished -- see \citet{boyce01, lang03, wolfinger13}, and \citet{davies04} for the three times deeper VIRGO\HI\ survey of the Virgo Cluster. Furthermore for the 2MASS Tully-Fisher Survey (2MTF) \HI\ data have been published for 1497 targeted galaxies, of which 878 were detected. The galaxy catalogs are \citet{masters08, hong13, masters14b}. These were obtained with the Green Bank Telescope (GBT) and at Parkes with HIPASS; see Sect.~\ref{results} for a comparison between results obtained with these telescopes and the NRT and NIBLES. Here, we present the \HI\ survey undertaken at \nan, and limit ourselves to a short discussion of the results. In future papers in this series, we will present the results of our deeper Arecibo \HI\ observations (Butcher et al. 2016a, Paper II, submitted to A\&A), the bivariate luminosity function and \HI\ mass function (Butcher et al., in prep.), stacking of \HI\ spectra of undetected sources (Healy et al., in prep.) and further analyses of the sample. The NIBLES \HI\ data are also used for a comparison with local galaxies with extremely high specific Star Formation Rates (\citealt{lehnert16} and Lehnert et al., in prep.) In Sect.~\ref{sample} we describe the selection of the observed sample of galaxies, in Sect.~\ref{observations} the observations and data reduction, in Sect.~\ref{results} the results, including a summary of the problems encountered with various SDSS Data Releases (DRs), and in Sect.~\ref{discussion} we present a first, brief discussion.

\label{discussion} % Further discussion and analysis of the data presented here will be given in future papers in this series (e.g., Paper II; Butcher et al., in prep.; Healy et al., in prep.). Our goal was to observe a total of 3000 galaxies in the local Universe, distributed as uniformly as possible over the full range in absolute $z$-band magnitude detected in the SDSS, without selection on color. The final \Mz\ distribution is shown in Fig.~\ref{fig:detstatMr}, which contains all sources observed and detected, both clearly and marginally. For some galaxies the redshift and/or photometry changed as a result of moving from the SDSS DR5 data used to select them to the DR9 data used for further analysis (see Sect.~\ref{results}). \begin{figure} % \centering \includegraphics[width=9cm]{NIBLES_paperI_fig13_lo.jpg} \caption{Integrated $g$-$z$ color, in mag, as a function of absolute $z$-band magnitude, \Mz. All data were corrected for Galactic extinction following \citet{schlegel98}. Black dots represent clear NIBLES detections, gray dots marginal detections, and red triangles estimated non-detections. } \label{fig:g-zMzNIBLESALFA} \end{figure} % We observed 130-165 sources per half-magnitude bin over the absolute magnitude range $-23 < M_{\rm z} < -16.5$ mag (the target number was 150), and observed all objects known from the DR5 in the least-filled bins (see Fig.~\ref{fig:detstatMr}). For \Mz$>-21$ mag the overall detection rate is high, at about the 75\% level, and rather constant, but it starts to decline steadily for more luminous galaxies, towards the 30\% level at $-$23 mag (see also Fig.~\ref{fig:g-zMzNIBLESALFA}). Shown in Fig.~\ref{fig:FHIW50NIBELSALFA} is the integrated \HI\ line flux, \FHI, as a function of \Wfifty\ line width for the 1733 clear detections and 137 marginal detections, together with the line indicating the flux expected for a flat-topped 3$\sigma$ detection with a 2 mJy $rms$ noise level. For comparison, the same data are shown for the 11,941 ALFALFA high-quality detections (their Category 1) and 3100 marginal detections (their Category 2) from \citet{haynes11}. As noted in Sect.~\ref{observations}, the $rms$ noise level is not constant for the entire NIBLES sample (see Fig.~\ref{fig:rmsdist}), with more observing time spent on weak marginal detections and non-detections, telescope time permitting; the mean $rms$ for the sample is 2.1 mJy. The comparable overall $rms$ noise level of the NIBLES and ALFALFA surveys is reflected in their similar \FHI-\Wfifty\ distribution (Fig.~\ref{fig:FHIW50NIBELSALFA}), where also the effect of longer integrations for weak NIBLES marginal detections can be noted. The higher velocity resolution of the ALFALFA data, 10 \kms\ compared to 18 \kms\ for NIBLES, has resulted in a number of very narrow ALFALFA detections, more extreme than found by us. The distribution of the total \HI\ masses as a function of radial velocity is shown in Figs.~\ref{fig:MHIVNIBLES} and \ref{fig:MHIVNIBLESALFA}. The former shows both the clear and the marginal NIBLES detections as well as the estimated upper limits of the non-detections, whereas the latter shows the clear and the marginal detections of both NIBLES and ALFALFA. Excluded in both figures were the NIBLES sources which are definitely or probably confused by another galaxy within the telescope beam (flags $C1$ and $C2$ in the tables). For the sake of clarity, the velocity range is only plotted to the NIBLES limit of 12,000 \kms, whereas ALFALFA detections continue out to 18,000 \kms. Figure~\ref{fig:MHIVNIBLES} shows that the estimated upper limits to the \HI\ masses of NIBLES non-detections are quite conservative, as they are based on the upper envelope of the \Wtwenty\ line widths measured as a function of $r$-band luminosity (see Sect.~\ref{results}). At the highest velocities (V$>$8000 \kms), where in general the most luminous sources are located with the greatest expected line widths, the upper limits even tend to be higher than the NIBLES detections at the same redshift. As can be seen in ~\ref{fig:MHIVNIBLESALFA}, the NIBLES SDSS sources were selected at the lowest velocities in their \Mz\ bins, as far as practicable, whereas ALFALFA is a blind \HI\ survey, bound to detect high numbers of relatively high \HI\ mass objects at larger distances than the SDSS sources selected for NIBLES. The ALFALFA line fluxes from the $\alpha$.40 catalog are on average 1.45 times higher than our \nan\ values due to a flux calibration difference (see Sect.~\ref{comparison}), which corresponds to a difference of 0.16~dex in log(\MHI) -- see the vertical arrow in the plot. The integrated $g$-$z$ color as a function of absolute $z$-band magnitude is shown in ~\ref{fig:g-zMzNIBLESALFA} with different symbols for NIBLES clear detections, marginal detections, and non-detections. NIBLES sources were selected on \Mz, irrespective of color. The distribution shows the well-known ``red sequence'' and ``blue cloud'' loci, with a mixture of detections and non-detections in both. The highest concentration of non-detections occurs among the most luminous red systems, but these are not all elliptical systems and \HI\ detections do occur. Among the low-luminosity systems (at log(\Lr) 7.8-8.5) the non-detected dwarfs are predominantly red. We will study the underlying \HI\ properties of undetected galaxies using four times higher sensitivity Arecibo observations in Paper II. \subsection{The \MHIMstar\ -- \Mstar\ relationship, including \HIit\ non-detections} \label{MHIMstar}% The ratio of \HI\ and stellar masses, \MHIMstar, as a function of total stellar mass is shown in Fig.~\ref{fig:MHIMstar} for NIBLES detections, marginals, and estimated upper limits for non-detections. We did not use those sources we flagged as either $C1/C2$ (confused by another galaxy in the NRT beam), $R$ (resolved by the NRT beam) or $U$ (significant SDSS flux missing), as their \MHIMstar\ ratios will be either under- or overestimated by unknown amounts (see Sect.~\ref{results}). Overlaid on this plot are high-quality ALFALFA detections, and the mean relationship for four literature reference samples of local \HI-detected galaxies (pink line) from \citet{papastergis12}. We first explore the differences between the plotted properties for the various samples and their uncertainties, both relative and systematic. Like for the NIBLES data, the total stellar masses used here for the ALFALFA detections were taken by us from the SDSS ``added-value" MPA/JHU catalogs (see Sect.~\ref{results}). To estimate the the uncertainty in the stellar masses of individual galaxies we examined the +1 $\sigma$, -1 $\sigma$ and median mass estimates, and found a typical (mean) relative uncertainty of about 20\%. Matching positions with those of the 15,598 $\alpha$.40 catalog \HI\ detections resulted in 2500 matches. A caveat in the comparison between our MPA/JHU catalogs' total stellar masses for NIBLES and ALFALFA galaxies and those used for the reference samples in \citet{papastergis12} is that the latter used a somewhat different way to estimate stellar masses, see \citet{huang12b} for details. We do not have a simple way of estimating the systematic uncertainties between the two stellar mass estimate methods as stellar masses for individual galaxies are not given in \citet{papastergis12}. We will therefore ignore this uncertainty, but we note that in our experience different mass estimates usually agree within about 0.3~dex \citep[e.g.,][]{drory04, moustakas13, sorba15}. For the NIBLES \HI\ masses, we estimate a typical relative uncertainty, due to variations within the telescope system, of about 15\% and a systematic uncertainty of about 10\%, after comparison with flux scales of other telescopes (see Sect.~\ref{comparison}). Even when correcting the log(\MHI) of the ALFALFA $\alpha$.40 catalog detections by $-$0.16~dex, they lie on, and above, the upper envelope of the NIBLES detections. \begin{figure*} % \centering \includegraphics[width=18cm]{NIBLES_paperI_fig14_lo.jpg} \caption{The ratio of total \HI\ and stellar masses, \MHIMstar, as a function of total stellar mass, \Mstar\ (in \Msun). Black dots represent clear NIBLES detections, gray dots marginal detections, and red open triangles estimated upper limits for non-detections, whereas green dots represent the clear ALFALFA detections from \citet{haynes11}. All total stellar masses used for the NIBLES and ALFALFA galaxies were taken from the SDSS ``added-value" MPA/JHU catalogs (see Sect.~\ref{results}). The \HI\ masses of the ALFALFA detections were calculated in the same way as for the NIBLES sources, using simply a distance of $D$ = $V$/H$_0$, where the adopted Hubble constant is $H_0$ = 70 \kms\ Mpc$^{-1}$. The green vertical arrow on the right indicates the mean flux scale difference of 0.16~dex in log(\MHI) between the higher ALFALFA $\alpha$.40 catalog \HI\ line fluxes and the NIBLES values (see Sect.~\ref{comparison}). For all fits made to the NIBLES data, we excluded sources that are either resolved, confused or missing significant SDSS flux. The dark green line shows a linear regression fit made only to the NIBLES detections. The dark blue line shows a linear regression fit to all NIBLES data including our rather conservative estimates for \HI\ mass upper limits based on the upper envelope in the \Wtwenty-\Lr\ relationship (Fig.~\ref{fig:W20Lr}), whereas the light blue line shows a fit using low upper limit estimates based on the mean \Wtwenty-\Lr\ relationship (see text). The pink line shows the mean relationship derived for four literature reference samples of \HI-detected local galaxies from \citet{papastergis12}, which is based on a stellar mass estimation method different from the one used in the MPA/JHU catalogs (see Sect.~\label{MHIMstar}). For the NIBLES detections, we indicate the typical uncertainty in both quantities by the black cross on the right hand side of the plot (see text for details). } \label{fig:MHIMstar} \end{figure*} % The mean relationship found by Papastergis et al. for the ensemble of four literature samples they use to evaluate the gas-to-stellar mass ratio of galaxies as a function of \Mstar\ (pink line) is log(\MHIMstar) = $-$0.43 log(\Mstar) + 3.75. The latter reference samples only contain objects selected based on previously known \HI\ detections, and do not include any \HI\ non-detections. Among the total of about 1000 galaxies, the largest sample used \citep{zhang09} is based on 721 \HI\ detections from HyperLeda, two others are based on Westerbork radio synthesis observations \citep{swaters02, noordermeer05} of 239 galaxies selected to have a strong enough \HI\ line to enable imaging and the smallest \citep{garnett02} contains 42 \HI-detected objects. Their \HI\ masses as used by Papastergis et al. are based on the originally published values, and are independent of subsequent ALFALFA measurements. Papastergis et al. also estimated maximum upper limits to \MHIMstar\ ratios for galaxies in their selected sky regions covered by ALFALFA which were not detected by ALFALFA, using \HI\ mass upper limits based on the 25\% completeness limit of the $\alpha$.40 catalog as a function of \HI\ line width. These ratios also lie systematically above the relationship for the four literature samples. In order to keep Fig.~\ref{fig:MHIMstar} readable, we did not plot the uncertainties for all individual galaxies but instead indicated the typical uncertainty for both the stellar mass and the \HI-to-stellar mass ratio of the NIBLES detections. For the \HI\ mass, we added the typical relative and systematic uncertainties in quadrature. A second, though minor, caveat in comparing samples shown in Fig.~\ref{fig:MHIMstar} concerns the distance scales used. For NIBLES we used a pure Hubble-flow method and heliocentric velocities, whereas for published ALFALFA detections (including those used in \citealt{papastergis12}) a correction for peculiar motions was applied for $V_{\rm CMB}$$<$6000 \kms\ (see \citealt{haynes11} for details). The ALFALFA \HI\ masses shown in Fig.~\ref{fig:MHIMstar} were all calculated by us using the same method as for NIBLES. For the four reference samples used in \citet{papastergis12}, no individual distances are given, nor in the paper that comprises the bulk of those data, \citet{zhang09}. From the fact that the latter are all SDSS galaxies with pre-ALFALFA \HI\ detections listed in HyperLeda, we deduce they are relatively nearby objects with a mean velocity of a few thousand \kms. At such velocities, the ALFALFA distances are only about 7\% higher than the NIBLES values, corresponding to +0.03~dex in log(\MHI) -- small compared to the other typical uncertainties of individual galaxy data as shown in Fig.~\ref{fig:MHIMstar}. A linear regression fit made only to the NIBLES detections (excluding resolved and confused sources) results in a mean relationship (dark green line in Fig.~\ref{fig:MHIMstar}) of log(\MHIMstar) = $-$0.54 log(\Mstar) + 4.70. We now want to examine the impact of \HI\ non-detections, of which the estimated upper limits to their total \HI\ masses are routinely ignored, in the study of this relationship for our optically-selected NIBLES sample. To this purpose, we fitted our data, including \MHI\ upper limits, with three estimators of the slope and intercept, from the STSDAS statistics package \footnote{http://stsdas.stsci.edu/cgi-bin/gethelp.cgi?statistics}. These were the Buckley-James, expectation-maximization (EM) algorithm, and Schmitt binning methods of linear regression. It is beyond the scope of the paper to discuss the differing nature of these methods \citep[see][for details]{feigelson85, isobe86}. At low stellar masses, log(\Mstar)$\approxlt$ 7.5, the relationship between \MHIMstar\ and \Mstar\ apparently becomes non-linear, and we therefore excluded galaxies below this limit from these linear fits. It is sufficient to say here, that all three methods gave similar slopes and intercepts within 0.1~dex in log(\MHIMstar). Using the Buckley-James method, we find log(\MHIMstar) = $-$0.59 log(\Mstar) + 5.05. Since we were concerned that our estimated 3$\sigma$ \MHI\ upper limits might be overly conservative owing to our choice of the largest observed \Wtwenty\ line widths as a function of \Lr\ (see Fig.~\ref{fig:W20Lr}), we also explored how adjusting our upper limit estimates would influence the fits. To this end, we subtracted 0.25~dex from the upper envelope line shown in Fig.~\ref{fig:W20Lr}, which corresponds to the mean relationship between \Wtwenty\ and \Lr. In this case, we found log(\MHIMstar) = $-$0.62 log(\Mstar) + 5.20. We show these two fits, dark blue when using our conservative upper envelope limits and light blue based on the mean \Wtwenty-\Lr\ relationship, in Fig.~\ref{fig:MHIMstar}. The uncertainties in our fits are 0.03 in the slope and 0.3 in the zero-point. Although \citet{papastergis12} do not provide an estimate of the uncertainties in their fit, the spread of the data points in their Fig. 19 indicates a standard deviation of order $\pm$0.2 dex in log(\MHIMstar) around the mean relationship. Overall, our regression fits to all NIBLES data are not very dependent on the choice of line widths for \HI\ mass upper limit estimates based on our \nan\ data, which on average causes a difference of 0.25 dex in log(\MHI); the difference in log(\MHIMstar) between using the conservative, upper envelope widths and the mean widths amounts to only 0.03~dex at log(\Mstar) = 7.5 and goes up to 0.12~dex at 11.5. About one quarter of the NIBLES galaxies were not detected in \HI, and most of these have the highest stellar masses. Therefore, the effect of using lower \MHI\ estimates for non-detections will have the greatest effect at the high mass end. On the other hand, the difference between the Papastergis et al. \HI-detected reference samples and the NIBLES fit using conservative upper limits amounts to $-$0.08~dex at log(\Mstar) = 7.5 and goes up to 0.57~dex at log(\Mstar) = 11.5. However, keep in mind there may be a systematic difference of up to 0.3~dex between the two methods used for total stellar mass estimates. Our follow-up Arecibo detections of 72 \nan\ non-detections show that they lie mainly among the lower envelope of the \nan\ detections and marginals in Fig.~\ref{fig:MHIMstar}, indicating that \HI\ observations sufficiently sensitive to detect all our targets would significantly lower the mean \MHIMstar\ ratio over the entire range of \Mstar\ for our NIBLES sample of optically selected local galaxies. We will discuss this in further detail in future papers in this series.

1607.02787

1607

1607.03674_arXiv.txt

{ We analyze new high spatial resolution ($\sim$60\,pc) ALMA CO(2--1) observations of the isolated luminous infrared galaxy ESO~320-G030 ($d=48$\,Mpc) in combination with ancillary \textit{HST} optical and near-IR imaging as well as VLT\slash SINFONI near-IR integral field spectroscopy. We detect a high-velocity ($\sim$450\,km\,s$^{-1}$) spatially resolved (size$\sim$2.5\,kpc; dynamical time $\sim$3\,Myr) massive ($\sim$10$^7$\,\Msun; $\dot{M}\sim$2--8\,\Msun\,yr$^{-1}$) molecular outflow originated in the central $\sim$250\,pc. We observe a clumpy structure in the outflowing cold molecular gas with clump sizes between 60 and 150\,pc and masses between 10$^{5.5}$ and 10$^{6.4}$\,\Msun. The mass of the clumps decreases with increasing distance, while the velocity is approximately constant. Therefore, both the momentum and kinetic energy of the clumps decrease outwards. In the innermost ($\sim$100\,pc) part of the outflow, we measure a hot-to-cold molecular gas ratio of 7$\times$10$^{-5}$, which is similar to that measured in other resolved molecular outflows. We do not find evidence of an ionized phase in this outflow. The nuclear IR and radio properties are compatible with strong and highly obscured star-formation ($A_{\rm k}\sim4.6$\,mag; ${\rm SFR}\sim15$\,\Msun\,yr$^{-1}$). We do not find any evidence for the presence of an active galactic nucleus. We estimate that supernova explosions in the nuclear starburst ($\nu_{\rm SN}\sim0.2$\,yr$^{-1}$) can power the observed molecular outflow. The kinetic energy and radial momentum of the cold molecular phase of the outflow correspond to about 2\% and 20\%, respectively, of the supernovae output. The cold molecular outflow velocity is lower than the escape velocity, so the gas will likely return to the galaxy disk. The mass loading factor is $\sim$0.1--0.5, so the negative feedback due to this star-formation powered molecular outflow is probably limited. }

\label{s:intro} Theoretical models predict that massive gas outflows, driven by starbursts or active galactic nuclei (AGN), are fundamental actors in shaping the observed properties of galaxies (e.g., galaxy mass function, mass-metallicity relation). This is because massive outflows can regulate both the accretion rate of the central supermassive black hole and the star-formation (SF) activity, but also because they can redistribute dust and metals over kpc scales (e.g., \citealt{Veilleux2005, Narayanan2008, Hopkins2012}). Gas outflows have a multiphase (ionized, neutral atomic, and molecular) structure \citep{Veilleux2005, Hopkins2012}. The ionized and neutral atomic phases have been studied over the past 25 years using optical and ultraviolet spectral features (e.g., \citealt{Heckman1990, Veilleux1995, Rupke2008, Arribas2014, Heckman2015, Cazzoli2016}). Just recently, thanks to improved capabilities of millimeter observatories, many works have focused on the molecular phase of the outflow, which dominates the outflowing gas mass, energy, and momentum in most cases (e.g., \citealt{Feruglio2010, Tsai2012, Bolatto2013ngc253, Cicone2014, GarciaBurillo2014, GarciaBurillo2015}). These studies show that massive molecular outflows are ubiquitous in AGN and starbursts and that they might have an important impact in the evolution of their host galaxies. Multi-transition and multi-specie studies of the molecular phase of the outflow shed some light on the poorly known physical and chemical properties of this phase. For instance, the isotopic O abundance measured using the far-IR OH absorption in the outflow of Mrk~231 suggests that the outflowing gas has been processed by advanced starbursts \citep{Fischer2010}. In addition, the HCN, HNC, and HCO$^+$ emissions indicate that the molecular gas in that outflow can be chemically differentiated while being compressed and fragmented by shocks \citep{Aalto2012, Lindberg2016}. Actually, the analysis of cold and hot molecular gas tracers (e.g., CO and near-infrared H$_2$ transitions) in various objects indicates that the outflowing molecular gas is continuously heated by shocks \citep{Dasyra2014, Emonts2014}. {Similarly, spatially resolved observations of outflows at sub-kpc scales (e.g., \citealt{Bolatto2013ngc253, Emonts2014, Sakamoto2014, GarciaBurillo2014, Salak2016}) are essential to study the outflowing molecular gas. The clumpy distribution of the molecular gas in the outflow, as shown in this work, and the physical properties of these clumps provide a benchmark for models and simulations to establish how the molecular gas evolves in the outflow (e.g., \citealt{Nath2009, Zubovas2014b}) and also to better constrain the global effect of the outflow in the host galaxy. } Many of these {outflow} studies analyze local luminous and ultraluminous infrared galaxies (LIRGs and ULIRGs respectively) because they are the most extreme examples of starbursts (and AGN) in the local Universe. LIRGs have infrared (IR) luminosities $>$10$^{11}$\,\Lsun, while ULIRGs have $L_{\rm IR}>10^{12}$\,\Lsun. These IR luminosities, if produced solely by star-formation, are equivalent to star-formation rates (SFR) of $>$15 and $>$150\,\Msun\,yr$^{-1}$ , respectively {(see Table 1 of \citealt{Kennicutt2012}).} These local U\slash LIRGs are considered as local counterparts of high-$z$ starburst with similar or higher IR luminosities (e.g., \citealt{Muzzin2010}). Therefore, they are perfect targets for detailed studies at high spatial resolution of the processes, for example outflows, taking place in their high-$z$ counterparts. In this paper, we analyze new Atacama Large Millimeter/submillimeter Array (ALMA) CO(2--1) high spatial resolution ($\sim$60\,pc) observations of the local ($d=48$\,Mpc; {scale of }240\,pc\,arcsec$^{-1}$) LIRG ESO~320-G030 ($\log L_{\rm IR}\slash \Lsun = 11.3$; also known as IRAS~F11506-3851). We combine the CO(2--1) data (cold molecular gas tracer) with available near-IR integral field spectroscopy (IFS) to trace the hot molecular gas as well as the ionized gas \citep{Piqueras2012}. This galaxy is an isolated spiral galaxy with an ordered velocity field \citep{Bellocchi2013} hosting a strong starburst \citep{AAH06s}. Its nuclear activity is classified as \ion{H}{ii} from optical spectroscopy \citep{vandenBroek1991, Pereira2011} and there is no evidence of an AGN in this galaxy based on its mid-IR and X-ray emissions \citep{Pereira2010c, Pereira2011}. It also host an OH megamaser \citep{Norris1986}, which are associated with compact starbursts most of the times \citep{Baan2006, Zhang2014}. In addition, a massive outflow of neutral atomic gas is already detected in this object using optical IFS \citep{Cazzoli2014, Cazzoli2016}. This paper is organized as follows: we describe the observations and data reduction in Section \ref{s:data}. The analysis of the gas and stellar morphology and kinematics are presented in Sections \ref{s:morphology} and \ref{s:vfield}. In Section \ref{s:outflow}, we describe the structure and physical properties of the resolved molecular outflow. In Section \ref{s:nature}, we investigate the nature of the compact central source of this galaxy. The main results are discussed in Section \ref{s:discussion} and, finally, in Section \ref{s:conclusions}, we summarize the main findings of the paper.

\label{s:conclusions} We present high spatial resolution ($\sim$60\,pc) ALMA CO(2--1) observations of the local spiral LIRG ESO~320-G030. We study the morphology and kinematics of the cold molecular gas traced by the CO(2--1) emission and combine these data with ancillary \textit{HST} optical and near-IR imaging as well as VLT\slash SINFONI near-IR integral field spectroscopy. The main goals of this work are: characterize the resolved massive molecular outflow detected in this object and establish the nature of the extremely obscured nucleus which produces the massive molecular outflow. The main results are the following: \begin{enumerate} \item The global kinematics is well represented by a regular rotating disk, although we find non-circular molecular gas motions related to the secondary bar. This can be the signature of inflowing gas motions, which might explain the extreme concentration of molecular gas in the central 500\,pc of ESO~320-G030 (60\% of the total CO(2--1) emission and 10$^{4.4}$\,\Msun\,pc$^{-2}$). \item We spatially resolve a high velocity ($\sim$450\,km\,s$^{-1}$) outflow containing 10$^{6.8}$\,\Msun\ of molecular gas (assuming the ULIRG conversion factor) originating in the central $\sim$250\,pc. The size of the outflow is $\sim\pm$1.2\,kpc, which corresponds to a dynamical time of $\sim$2.8\,Myr. The opening angle is $\sim$30\degree. \item We measure the properties of 7 clumps in the outflow. Their sizes are 60--150\,pc and they have molecular gas masses between 10$^{5.5}$ and 10$^{6.4}$\,\Msun (assuming an ULIRG-like conversion factor). The mass, kinetic energy, and momentum of the clumps decrease with increasing distances while the velocity is approximately constant. \item The hot molecular gas component of the outflow, as probed by the near-IR H$_2$ transitions, is detected in the innermost ($\sim$100\,pc) part of the outflow with a hot-to-cold molecular gas ratio of 7$\times$10$^{-5}$. This ratio is similar to that measured by \citet{Emonts2014} in another resolved molecular outflow. \item We find that the nuclear IR and radio emission of the nucleus ($d\sim250$\,pc) are compatible with highly obscured intense SF ($A_{\rm k}\sim$\,4.6\,mag; SFR\,$\sim$\,15\,\Msun\,yr$^{-1}$). No evidence for the presence of an AGN is found. The outflow rate and loading factor are $\dot{M}_{\rm out}=$2--8\,\Msun\,yr$^{-1}$ and $\sim$0.1--0.5, respectively, and depending on the CO conversion factor assumed. This low loading factor indicates that star-formation quenching due to the molecular outflow is not efficient in this object. \item We find that SN explosions in the nuclear starburst ($\nu_{\rm SN}\sim$0.2\,yr$^{-1}$) can power the observed molecular outflow. The kinetic energy and radial momentum of the outflow represent $\sim$2\% and 20\%, respectively, of the SNe output. \item The velocity of the outflowing clumps is lower than the escape velocity, so it is likely that most of this molecular gas will return to the disk after several Myr. This is compatible with the thick neutral atomic gas disk found in this object. \end{enumerate} {The origin of this outflowing molecular gas (either formed in overdensities in the outflowing ionized\slash neutral gas or dragged from the nuclear ISM) is not clear from the available data. New high-spatial resolution observations of the different phases of this outflow will help to establish the formation and evolution of the observed molecular gas. }

1607.03674

1607

1607.04262_arXiv.txt

We report on the discovery of a hydrogen-deficient compact binary (CXOGBS J175107.6-294037) belonging to the AM CVn class in the Galactic Bulge Survey. Deep archival X-ray observations constrain the X-ray positional uncertainty of the source to 0\farcs57, and allow us to uniquely identify the optical and UV counterpart. Optical spectroscopic observations reveal the presence of broad, shallow He \textsc{i} absorption lines while no sign of hydrogen is present, consistent with a high state system. We present the optical lightcurve from Optical Gravitational Lensing Experiment monitoring, spanning 15 years. It shows no evidence for outbursts; variability is present at the 0.2 mag level on timescales ranging from hours to weeks. A modulation on a timescale of years is also observed. A Lomb-Scargle analysis of the optical lightcurves shows two significant periodicities at 22.90 and 23.22 min. Although the physical interpretation is uncertain, such timescales are in line with expectations for the orbital and superhump periods. We estimate the distance to the source to be between 0.5\,--\,1.1 kpc. Spectroscopic follow-up observations are required to establish the orbital period, and to determine whether this source can serve as a verification binary for the \textit{eLISA} gravitational wave mission.

\label{sec:introduction} AM Canum Venaticorum (AM CVn) systems are a class of binary stars, consisting of a white dwarf (WD) accreting H-deficient material from a low mass companion. These binaries are in a very compact configuration with observed orbital periods ranging from 5 to 65 min, suggesting the companion is a (semi-) degenerate star (see e.g. \citeauthor{Solheim2010} \citeyear{Solheim2010} for a recent review). The class of AM CVn stars can be divided into four groups, depending on the evolutionary stage of the binary: \begin{enumerate} \item P$_{\text{orb}} \lesssim$ 10 min: direct impact systems without a disc \item 10 min $\lesssim$ P$_{\text{orb}} \lesssim$ 20 min: stable disc in a persistent high state \item 20 min $\lesssim$ P$_{\text{orb}} \lesssim$ 40 min: variable disc with outbursts \item 40 min $\lesssim$ P$_{\text{orb}} \lesssim$ 65 min: stable disc in quiescence \end{enumerate} The orbital evolution of these short period systems is expected to be dominated by gravitational wave radiation (GWR; \citeauthor{Tutukov1979} \citeyear{Tutukov1979}). In that case, the mass transfer rate is set by angular momentum loss due to GWR, and because the latter is a steep function of the orbital period, this implies a strong decrease of $\dot{M}$ with increasing P$_{orb}$. This provides a natural explanation for the observed behaviour. With decreasing $\dot{M}$, the evolutionary timescale will become longer. Most sources belong to classes (iii) and (iv), while only two direct impact systems and four persistent high state sources are currently known. The transition between groups (ii) and (iii) is thought to occur around P$_{\text{orb}}$\,$\sim$\,20 min, but the exact moment at which the transition occurs is currently not well constrained. \citet{Tsugawa1998} present a disc instability model for He-rich accretion discs in AM CVns (see also \citeauthor{Kotko2012} \citeyear{Kotko2012}), in analogy with the H-rich discs in cataclysmic variables \citep{Smak1982}. They showed that, depending on the accretion rate and accretor mass (which set the temperature at the outer and inner edges of the disc, respectively), three distinct situations can occur: if the mass transfer rate is very high, the material in the disc will be fully ionized and the disc is in a hot, stable state. If the mass transfer rate is very low, no ionized material is present in the disc and the system is in a cool, stable state. For mass transfer rates in between these two limiting cases, the temperature will increase as more material is transferred from the donor to the disc. When the temperature surpasses the ionization temperature of He, this leads to an increase in the viscosity in the disc. When this thermal instability occurs, the material in the disc is rapidly accreted onto the WD and an outburst is observed. Other system properties such as the chemical composition will influence whether the disc remains stable or goes into outburst \citep{Kotko2012}. In this letter we present the identification of CXOGBS J175107.6-294037 (also known and from here-on referred to as CX361) as a persistent high state AM CVn binary.

\label{sec:discussion} We have discovered a persistent X-ray source, CX361, in the Galactic Bulge Survey. We obtained optical spectroscopy, presented in Figure \ref{fig:opticalspectra}. The spectrum shows shallow, broad and asymmetric He\,\textsc{i} absorption lines, lacking signatures of hydrogen. This is typical for AM CVn binaries in a high accretion state. The spectrum looks very similar to high state systems such as AM CVn itself \citep{Roelofs2006} and HP Lib \citep{Roelofs2007}. Small amplitude variability is observed in the optical lightcurve on timescales of hours, days and weeks at a level $\sim$0.2 mag (Figure \ref{fig:oglelc}). This is consistent with an origin in an accretion disc, and similar to the photometric properties of high state AM CVn binaries \citep{Skillman1999, Patterson2002}. Our photometric observations also show that no outbursts have been detected in the last 15 years. Because the typical outbursts of an AM CVn system with P$_{\text{orb}} \sim$ 23 min (see below for an explanation) last 15 days, recur every 40 days and have an amplitude of $\sim$3 mag \citep{Levitan2015}, these would have been detected in our observations. Given that the spectrum indicates the system is in a high state, the lack of large amplitude variability suggests that the disc is persistently in the high state. On top of the flickering, there is evidence for a long term modulation of the brightness on a timescale of years. This could indicate deviations from the secular mass transfer rate set by the GWR, or due to variable irradiation of the donor by disc precession and/or warping \citep{Kotko2012}. \begin{figure} \includegraphics[height=4.4cm, keepaspectratio]{oglelc.pdf} \caption{Long term \textit{I}-band lightcurve of CX361 from OGLE III (t\,$\leq$\,5000) and IV (t\,$\geq$\,5000). Variability is present at the 0.2 mag level on hours, days and weeks timescales. There is also evidence for a long term modulation of the brightness. The dashed lines mark the epochs of the SOAR (left) and VIMOS (right) spectra.} \label{fig:oglelc} \end{figure} \begin{figure} \includegraphics[height=8.5cm, width=0.5\textwidth]{psd_orbital+superhump.pdf} \caption{Top: LS periodogram of the detrended OGLE IV lightcurve. The arrows indicate two periodicities 22.90 and 23.22 min. The 5$\sigma$ detection level is marked by a dashed line. Middle: OGLE lightcurve, binned and folded on the 22.90 min period. Bottom: detrended OGLE IV lightcurve, binned and folded on the 23.22 min period. The zero phase is arbitrary.} \label{fig:psd} \end{figure} A Lomb-Scargle analysis of the optical lightcurves shows evidence for two significant periodicities at P$_1$\,=\,22.90\,$\pm$\,0.01 and P$_2$\,=\,23.22\,$\pm$\,0.01 min (Figure \ref{fig:psd}). Their 1 day aliases are clearly visible in the periodogram. Typically, two strong periodicities are discovered in high state AM CVn sources: one at the orbital period and one at the superhump period. These periods are closely related, as the superhump period is likely caused by periodically modulated dissipation of the kinetic energy of the accretion stream due to variable irradiation \citep{Smak2009, Smak2013}. The photometric period may be related to the orbital period \citep{Levitan2011}, but this does not hold for all systems (see e.g. \citeauthor{Morales2003} \citeyear{Morales2003}). We show the binned lightcurve of the OGLE data folded on P$_1$ and P$_2$ in the middle and bottom panels of Figure \ref{fig:psd}, respectively. For P$_1$, the amplitude of variability over one cycle is $\sim$\,0.03 mag with a typical scatter of $\sim$\,0.01 mag. For P$_2$, the amplitude is 0.02 mag with 0.01 mag scatter. Comparing the morphology of these phase folded, binned lightcurves to those of HP Lib \citep{Patterson2002} and AM CVn \citep{Skillman1999}, we conclude that P$_1$ is likely the orbital period and P$_2$ the superhump period. We note that folding the entire OGLE lightcurve on P$_2$ does not result in a clean periodicity. For that reason, we show the detrended OGLE IV data folded on P$_2$ instead. This could suggest that there are slight variations in the superhump period over time. The difference between P$_1$ and P$_2$ is 19~s. Assuming that P$_1$\,=\,P$_{\text{orb}}$ and P$_2$\,=\,P$_{\text{sh}}$, this leads to a period excess ($\epsilon$\,=\,$\frac{P_{\text{sh}}\,-\,P_{\text{orb}}}{P_{\text{orb}}}$) of $\epsilon$\,=\,0.0138, similar to that of HP Lib ($\epsilon$\,=\,0.0148). This is consistent with the (tentative) relation between $\epsilon$ and P$_{\text{orb}}$ for systems with a dynamically determined orbital period (fig. 27 in \citeauthor{Solheim2010} \citeyear{Solheim2010}). We can infer a mass transfer rate of $\dot{M}$\,=\,10$^{-9}$~M$_{\odot}$~yr$^{-1}$ by using the relation between P$_{\text{orb}}$ and $\dot{M}$ found by \citet{Cannizzo2015}. Comparing this with the region of He disc stability (\citeauthor{Tsugawa1998} \citeyear{Tsugawa1998}; see also fig. 2 in \citeauthor{Deloye2005} \citeyear{Deloye2005}), we see that the system falls in the stable disc regime, consistent with our interpretation. We rule out a scenario with an ultra-compact X-ray binary (UCXB) in outburst based on the low observed $\frac{\text{F}_x}{\text{F}_{\text{opt}}}$ ratio. \citet{Jonker2011} calculate the typical X-ray fluxes and optical magnitudes for a population of AM CVns and UCXBs based on observed properties. Comparing our measurements to their populations (fig. 4 in \citeauthor{Jonker2011} \citeyear{Jonker2011}), we see that they are inconsistent with an UCXB scenario. In that case, the system should have at least an order of magnitude higher $\frac{\text{F}_x}{\text{F}_{\text{opt}}}$ ratio than observed. Phase-resolved spectroscopy is required to unambiguously determine the orbital period. If confirmed at 22.90 min, the orbital period would be similar to that of a confirmed outbursting system (22.5 min for PTF1 J191905.19+481506.2; \citeauthor{Levitan2014} \citeyear{Levitan2014}). This would provide direct observational evidence that in addition to P$_{\text{orb}}$, other factors play an important role in the disc stability of high state AM CVn systems. Comparing our values of $\dot{M}$ and P$_{\text{orb}}$ to fig. 2 of \citet{Deloye2007conf}, we conclude that our estimates are inconsistent with a zero-temperature WD but rather suggest a semi-degenerate donor star (either a high-entropy WD or He star). The implied high-entropy donor star may in part explain why the disc in this system is stable while PTF1 J1919, at similar P$_{\text{orb}}$, shows outbursts. Using the X-ray observations, we can obtain two distance estimates for the system. By using the relation between the hydrogen column density (N$_H$, determined from our fits to the X-ray spectra) and optical extinction found by \citet{Guver2009}, we estimate that the extinction in the \textit{V}-band is A$_V$\,=\,0.7 mag. Using the 3D reddening map by \citet{Green2015}, this corresponds to a distance of $\sim$ 500 pc. Alternatively, assuming that the X-ray luminosity is equal to that of HP Lib (1.4\,$\times$\,10$^{31}$~erg~s$^{-1}$; \citeauthor{Ramsay2006} \citeyear{Ramsay2006}) we estimate a distance to CX361 of 1.1 kpc. A short orbital period would make this source an excellent candidate to serve as a verification source for the future \textit{eLISA} space-based gravitational wave mission \citep{Amaro2012}. Assuming AM CVn-like parameters, but scaling for the tentative orbital period and distance of CX361 we can get a rough estimate of the GW strain (\textit{h}). Using \textit{h}\,=\,2.1$\times$10$^{-22}$ for AM CVn \citep{Roelofs2006}, we find \textit{h}\,$\sim$\,9\,$\times$\,10$^{-23}$ at 1.1 kpc, roughly a factor 2 above the confusion-limited Galactic background predicted by \citet{Nelemans2004}. A determination of system parameters such as an accurate distance and inclination are needed to confirm the expected GW strain caused by this source.

1607.04262

1607

1607.01392_arXiv.txt

The unprecedented extent of coverage provided by \Kep\ observations recently revealed outbursts in two hydrogen-atmosphere pulsating white dwarfs (DAVs) that cause hours-long increases in the overall mean flux of up to 14\%. We have identified two new outbursting pulsating white dwarfs in \Ktwo, bringing the total number of known outbursting white dwarfs to four. EPIC\,211629697, with \teff\ = $10{,}780 \pm 140$\,K and \logg{} = $7.94 \pm 0.08$, shows outbursts recurring on average every 5.0\,d, increasing the overall flux by up to 15\%. EPIC\,229227292, with \teff\ = $11{,}190 \pm 170$\,K and \logg\ = $8.02 \pm 0.05$, has outbursts that recur roughly every 2.4\,d with amplitudes up to 9\%. We establish that only the coolest pulsating white dwarfs within a small temperature range near the cool, red edge of the DAV instability strip exhibit these outbursts.

White dwarf stars are the remnant products of 97\% of Galactic stellar evolution. About 80\% of white dwarfs spectroscopically display atmospheres dominated by hydrogen \citep[DA;][]{Tremblay2008}. Convective driving \citep{Brickhill1991,Goldreich1999} of nonradial gravity-mode pulsations \citep{Robinson1982} in DA white dwarfs between 12,500~K $>$ \teff\ $>$ 10,600~K \citep[for typical \logg\ $\approx 8.0$;][]{Tremblay2015} causes these objects to appear photometrically variable. The frequencies of photometric variability are eigenfrequencies of these stars as physical systems, providing a powerful tool for studying their interior structures \citep[see reviews by][]{Winget2008,Fontaine2008,Althaus2010}. The \Kep\ spacecraft has provided unrivaled monitoring of pulsating white dwarfs, both in its original mission and during the two-reaction-wheel mission, \Ktwo\ \citep{Howell2014}. The first and longest-observed pulsating DA white dwarf (DAV) known to lie within the original mission field is KIC\,4552982 \citep[WD~J191643.83+393849.7;][]{Hermes2011}. This target was observed nearly continuously every minute for more than 1.5\,yr. Unexpectedly, these data revealed at least 178 brightness increases that recurred stochastically on an average timescale of 2.7\,d. The events increased the total flux output of the star by $2-17$\% and lasted $4-25$\,hr \citep{Bell2015}. \citet{Hermes2015} described a second DAV to display similar outburst behavior: PG\,1149+057, observed in \Ktwo\ Campaign 1. These outbursts caused the mean flux level to increase by up to 14\%, which would correspond to a nearly 750\,K global increase in the stellar effective temperature, with a recurrence timescale of roughly 8\,d and a median duration of 15\,hr. Mean pulsation frequencies and amplitudes were both observed to increase in this star during outbursts, and the combined flux enhancement from outbursts and high amplitude pulsations reached as high as 45\%. The outbursts affect the pulsation properties of PG\,1149+057, and \citet{Hermes2015} unambiguously ruled out a close companion or a line-of-sight contaminant as the source of this phenomenon. Spectroscopic effective temperatures place both of these white dwarfs very near to the empirical cool edge of the DAV instability strip---the boundary below which pulsations have not been detected in white dwarfs. While nonadiabatic pulsation codes successfully reproduce the observed hot edge of the DAV instability strip, they typically predict a cool edge thousands of Kelvin below what is observed \citep[e.g.,][]{VanGrootel2012}. The discovery of a new astrophysical phenomenon that operates precisely where our models are discrepant with observations suggests that the continued discovery and study of cool outbursting DAVs may inform fundamental improvements to the theory of stellar pulsations. In this paper we present the identification of two new outbursting DAVs that were observed by \Ktwo\, along with one candidate outburster. EPIC\,211629697 was observed at short cadence in \Ktwo\ Campaign 5 and EPIC\,229227292 in Campaign 6. Both stars are qualitatively similar in outburst and pulsational properties to the two previously published objects. We characterize these stars in Sections 2 and 3, respectively. Additionally, we inspect the light curves of the hundreds of other white dwarfs already observed by \Ktwo\ in Section 4, and describe a candidate single outburst in the long-cadence data of EPIC\,211891315. We summarize the current members of the outbursting class of DAV and discuss possible physical mechanisms and outburst selection effects in Section 5. \begin{figure*} \centering \includegraphics[width=0.99\columnwidth]{ktwo211629697_LC.pdf} \caption{\emph{Left:} The \Ktwo{} Campaign 5 light curve of EPIC\,211629697. The short-cadence data are displayed in black (during quiescence) and gray (during the 15 detected outbursts). The long-cadence data are shown in red. \emph{Right:} A detailed view of the outburst of median energy (see text). The units on the x-axes are the same in both panels. The scales of the y-axes are identical, with greater apparent scatter in the left panel due only to the overlap of points.} \label{fig:oDAV3lc} \end{figure*} \begin{figure} \centering \includegraphics[width=0.99\columnwidth]{ktwo211629697_FT.eps} \caption{Fourier transform of the entire \Ktwo\ light curve of EPIC\,211629697, including outbursts. The dashed line gives the 0.1\% False Alarm Probability (FAP) significance threshold for a single peak determined from bootstrapping (see text). The peak at 2053.514\,\muhz\ (486.97\,s) and 3 frequencies in the range 764--913\,\muhz\ (1095--1309\,s) reach amplitudes that exceed this significance threshold. We discard all low-frequency peaks below 100\,\muhz\ (see text).} \label{fig:oDAV3ft} \end{figure}

\label{sec:conc} The four confirmed members of the outbursting class of DAV have three distinct commonalities: (1) repeated outbursts, recurring on irregular intervals of order days and lasting for several hours; (2) effective temperatures that put them near the cool, red edge of the DAV instability strip; and (3) rich pulsation spectra dominated by low-frequency ($800-1400$\,s period) pulsations that are unstable in amplitude/frequency with at least one stable mode at significantly higher frequency ($350-515$\,s, and maybe as short as 290\,s), which in the first two cases appeared to be an $\ell=1$ from rotational splittings \citep{Bell2015,Hermes2015}. We summarize their main characteristics in Table~\ref{tab:summary}. \begin{deluxetable*}{c c c c c c c c c}[t] \tablecolumns{9} \tablecaption{Properties of Outbursting DAVs \label{tab:summary}} \tablehead{ \colhead{Name} & \colhead{$K_p$} & \colhead{\teff} & \colhead{\logg} & \colhead{$\tau_{recur}$} & \colhead{Med. Duration} & \colhead{Max. Flux} & \colhead{Max. Energy} & \colhead{Reference} \\ \colhead{} & \colhead{(mag)} & \colhead{(K)} & \colhead{(cgs)} & \colhead{(d)} & \colhead{(hr)} & \colhead{(\%)} & \colhead{(erg)} & \colhead{} } \startdata KIC\,4552982 & 17.9 & $10{,}860(120)$ & $8.16(0.06)$ & 2.7 & 9.6 & 17 & $2.1\times 10^{33}$ & \citet{Bell2015} \\ PG\,1149+057 & 15.0 & $11{,}060(170)$ & $8.06(0.05)$ & 8.0 & 15 & 45 & $1.2\times 10^{34}$ & \citet{Hermes2015} \\ EPIC\,211629697 & 18.4 & $10{,}570(120)$ & $7.92(0.07)$ & 5.0 & 16.3 & 15 & $1.8\times 10^{34}$ & This work \\ EPIC\,229227292 & 16.7 & $11{,}190(170)$ & $8.02(0.05)$ & 2.4 & 10.2 & 9 & $3.1\times 10^{33}$ & This work \enddata \end{deluxetable*} The discovery of repeated outbursts in four of the first 16 DAVs observed by the {\em Kepler} spacecraft indicates that this is not an incredibly rare phenomenon. However, it does beg the question of how outbursts have been missed during the first 45 years of studies of pulsating white dwarfs. In this context, the minimum outburst duration observed offers a clue: So far, every outburst lasts for more than several hours. Nearly all previous ground-based, time-series photometry of pulsating white dwarfs involves differential photometry: dividing the target by a (usually redder) comparison star to compensate for changing atmospheric conditions. Due to color-dependent extinction effects, nearly all groups have adopted a methodology of dividing out at least a second-order polynomial to normalize the light curves \citep[e.g.][]{Nather1990}. It is possible that outbursts were observed during previous ground-based studies of pulsating white dwarfs but were unintentionally de-trended from the data. Notably, the DBV (pulsating helium-atmosphere white dwarf) GD 358 underwent a large-scale brightening event in 1996, which may have been the first documented case of an outburst in a pulsating white dwarf \citep{Nitta1999,Montgomery2010}. The physical mechanism that causes outbursts remains an exciting open question. \citet{Hermes2015} suggested that, following the theoretical framework laid out by \citet{Wu2001}, the outbursts could be the result of nonlinear three-mode resonant coupling. In this model, energy is transferred from an observed, overstable parent mode to daughter modes via parametric resonance, one or both of which may be damped by turbulence in the convection zone and deposit their newfound energy there. All four of the outbursting white dwarfs have some of the longest pulsation periods observed in DAVs, excluding the extremely low-mass white dwarfs \citep{Hermes2013}. \citet{Wu2001} predicted that mode coupling would be most prevalent in the coolest white dwarfs with the longest-period pulsations, simply because there are more possible modes with which to couple. By inspecting the light curves of the more than 300 spectroscopically confirmed DA white dwarfs observed already by {\em K2}, we have shown that outbursts only occur in a narrow temperature range, between roughly $11{,}300$\,K and $10{,}600$\,K. This temperature range falls just hot of the empirical red edge of the DAV instability strip, below which pulsations are no longer observed. The red edge of the DAV instability strip has been notoriously difficult to predict from nonadiabatic pulsation codes, which suggest that white dwarfs should have observable pulsations down to at least 6000\,K \citep[e.g.][]{VanGrootel2012}. There have been two proposed mechanisms to bring the theoretical red edge in line with observations. \citet{Hansen1985} suggested that there is a critically maximal mode period, beyond which $g$-modes are no longer reflected off the outer mode cavity and thus evanesce. \citet{VanGrootel2013} showed that applying this critical mode period for $\ell=1$ modes to the thermal timescale at the base of the convection zone can successfully reproduce the empirical red edge of the DAV instability strip across a wide range of white dwarf masses. Additionally, a series of papers by Wu \& Goldreich proposed amplitude saturation mechanisms in the coolest DAVs from turbulent viscosity of the convection zone as well as resonant three-mode interactions as ways to cause a hotter red edge than nonadiabatic predictions \citep{Goldreich1999b,Wu2001}. If outbursts are indeed caused by nonlinear mode coupling, this suggests amplitude saturation as an important contributor to the cessation of observability of pulsations in the coolest DAVs. The measured properties of outbursts provide observational leverage for efforts to understand pulsational mode selection and driving, especially in the context of the few short-period modes that are selected in all four of the outbursting DAVs. Fortunately, DAV pulsations are extremely sensitive to structural changes in white dwarfs, and our understanding of outbursts will benefit from further asteroseismic analysis of these objects that will be the subject of future work. \Ktwo\ continues to obtain extensive space-based photometry on new fields roughly every three months, and we look forward to inspecting future data releases for additional instances of this exciting physical phenomenon.

1607.01392

1607

1607.06098_arXiv.txt

We search for a galaxy clustering bias due to a modulation of galaxy number with the baryon-dark matter relative velocity resulting from recombination-era physics. We find no detected signal and place the constraint $b_v < 0.01$ on the relative velocity bias for the CMASS galaxies. This bias is an important potential systematic of Baryon Acoustic Oscillation (BAO) method measurements of the cosmic distance scale using the 2-point clustering. Our limit on the relative velocity bias indicates a systematic shift of no more than $0.3\%$ rms in the distance scale inferred from the BAO feature in the BOSS 2-point clustering, well below the $1\%$ statistical error of this measurement. This constraint is the most stringent currently available and has important implications for the ability of upcoming large-scale structure surveys such as DESI to self-protect against the relative velocity as a possible systematic.

\label{sec:intro} Prior to decoupling at redshift $z\sim 1020$, baryons and dark matter behave differently because they experience different forces. The Universe is ionized, and the electrons are tightly coupled to the photons through Thomson scattering, while the protons follow the electrons under the Coulomb force. On scales within the sound horizon, the baryons are supported against gravitational infall by the photon pressure, which is large because the photons are an important component of the energy density and are relativistic. Consider the evolution of a point-like density perturbation in an otherwise homogeneous Universe. It will create a photon overpressure that launches a pulse of baryons and photons outwards (i.e. produce a Baryon Acoustic Oscillation (BAO)), and this pulse's front will be at the sound horizon (Sakharov 1966; Peebles \& Yu 1970; Sunyaev \& Zel'dovich 1970; Bond \& Efstathiou 1984, 1987; Holtzmann 1989; Hu \& Sugiyama 1996; Eisenstein \& Hu 1998; Eisenstein, Seo \& White 2007; Slepian \& Eisenstein 2016a). Baryons and photons farther away from the overdensity than the sound horizon will not yet know about the overpressure, and so they must infall under gravity. Meanwhile, the dark matter is insensitive to the photon pressure and infalls under gravity on all scales. As a result, when the photons release the baryons at decoupling, the dark matter within the sound horizon has a head start on infalling towards the initial density perturbation: there is a baryon-dark matter relative velocity on scales within the sound horizon. The magnitude of this relative velocity depends on the magnitude of the initial density perturbation. Therefore different regions of the Universe have different relative velocities, and these velocities are coherent on sound-horizon ($100\Mpch$) scales. The relative velocity effect was first calculated by Tseliakhovich \& Hirata (2010), and shortly after (Dalal, Pen \& Seljak 2010; Yoo, Dalal \& Seljak 2011) it was shown that this relative velocity can shift the BAO signal in the 2-point clustering if the late-time Luminous Red Galaxies (LRGs) used for these measurements have strong memories of their earliest progenitors. In particular, the relative velocity's root mean square value at $z\sim 50$, when the first galaxies are expected to form, is of order $10\%$ of the smallest $10^6\;M_{\odot}$ dark matter haloes' circular velocities or velocity dispersions. Thus small dark matter haloes living in a region of high relative velocity will find it difficult to capture baryons: the baryons' kinetic energy in the dark matter's rest frame is too large. The relative velocity can therefore induce an additional modulation of the clustering of these primordial galaxies on scales out to the BAO scale of $100\Mpch$. This modulation adds or subtracts from the primordial 2-Point Correlation Function (2PCF) within the BAO scale but not outside, and so can shift the BAO bump in or out in physical scale (this configuration space picture was devloped in Slepian \& Eisenstein 2015a, hereafter SE15a). If the late-time LRGs used for the BAO method at present have a strong memory of their early, small progenitors, the relative velocity can therefore bias the measured cosmic distance scale. Note that the relative velocity is fundamentally an effect set by the relativistic sound speed prior to decoupling; thus its large-scale coherence is unique and cannot be substantially modified by later-time feedback processes as or non-linear structure formation as they operate on far smaller scales. The most recent work on this bias has shown that even a small coupling of the relative velocity to late-time galaxy formation can induce a substantial shift in the distance scale (Blazek, McEwen \& Hirata 2016). Yoo, Dalal \& Seljak (2011) proposed that the bispectrum (Fourier space analog of the 3-Point Correlation Function (3PCF)) could be used to measure the relative velocity bias and then correct any effect in the 2PCF, but up to now this technique has not been used. Yoo \& Seljak (2013) used the power spectrum of galaxies to constrain the relative velocity bias in their bias model to be less than $0.033$; due to different normalization conventions this translates to a $b_v$ constraint in our bias model of $0.1$, as we further detail in \S\ref{sec:concs}. Beutler et al. (2016) compared different redshift slices within the Baryon Oscillation Spectroscopic Survey (BOSS) and WiggleZ to look for the relative velocity effect and found no evidence for it. Higher precision constraints are required: Blazek, McEwen \& Hirata (2016) show that even a relative velocity bias of order $4\%$ of the linear bias can cause a $1\%$ shift in the distance scale, comparable with the statistical errors on the lastest BOSS measurement (Cuesta et al. 2016; Gil-Mar\'in et al. 2016; Alam et al. 2016). Further, Dark Energy Spectroscopic Instrument (DESI; Levi et al. 2013), with first light in 2019, will improve on the BOSS error bars by roughly a factor of five, so even a relative velocity bias that is $1\%$ of the linear bias could systematically shift the distance scale comparably to DESI's statistical error bars. Given that any robustly detected deviation of the dark energy equation of state $w$ from $-1$ would have profound consequences for our understanding of dark energy, high precision on the relative velocity bias is required. In the present work, we use the 3PCF technique first proposed in Yoo, Dalal \& Seljak (2011) and developed further in SE15a to constrain the relative velocity bias to be $\leq 0.01$ for the SDSS BOSS DR12 CMASS galaxies (Eisenstein et al. 2011 for SDSS-III overview; Alam et al. 2015 for DR11 and DR12). This precision is sufficient to ensure that the cosmic distance scale measurement from BOSS will not be systematically biased. It also suggests our technique is powerful enough to allow DESI to avoid this bias. Our measurement is the most stringent constraint on the relative velocity bias available, and illustrates the power of the 3PCF for relative velocity constraints. The paper is laid out as follows. In \S\ref{sec:bias_model}, we present our galaxy bias model including relative velocity bias. \S\ref{sec:results} describes our relative velocity constraint, discusses the other bias parameters, and reports the cosmic distance scale measured with this bias model. We conclude in \S\ref{sec:concs}.

\label{sec:concs} We have shown that the 3PCF permits a $0.01$ precision measurement of the relative velocity bias, translating to about $0.5\%$ of the linear bias. We have shown three different estimates of this error bar that are all consistent with each other. For the data we find $b_v$ consistent with zero within our error bars. The constraint of Yoo \& Seljak (2013) from the power spectrum used $260,000$ SDSS DR11 galaxies to place the constraint $b_{v,{\rm YS}} < 0.033$. In their bias model, the relative velocity's square is normalized by its 1-D variance, $\sigma_{{\rm bc},\; 1-{\rm D}}^2$, which is $1/3$ the 3-D variance $\sigma_{\rm bc}^2$ used in our bias model (\ref{eqn:bias_model}). Holding the combination $b_v(\vbc^2/\sigma_{\rm bc}^2)$ constant, as this is what enters the bias models, the normalization difference means that their measured $b_v$ should be multiplied by a factor of $3$ to be compared with ours. Thus as we define $b_v$ the Yoo \& Seljak (2013) constraint is $b_v < 0.1$. The constraint $b_v<0.01$ of this work is a factor of ten tighter. Were our 3PCF technique equally good as the power spectrum analysis, we would expect a $0.58$ precision constraint (the precisions simply scale as $\sqrt{N_{\rm g}}$, with $N_{\rm g}$ the number of galaxies). Finding a $0.01$ constraint thus shows the superiority of the 3PCF for these measurements by roughly a factor of six. The relative velocity effect can bias the BAO scale measured in the 2PCF, and thus a tight constraint on $b_v$ is essential for present surveys such as BOSS and future efforts like DESI to remain unbiased. The constraint we find in this work is tight enough that the shift in $\alpha$ measured from the 2PCF will be less than $0.3\%$, using a recalculation of Blazek, McEwen \& Hirata (2016) Figure 2 for the appropriate linear bias and survey redshift for CMASSS to translate $b_v/b_1$ to a shift in $\alpha$. The BOSS survey can thus control $b_v$ to a level equivalent to $\sim 1/3$ of its BAO statistical precision. For surveys of larger volume at similar number density, this indicates that the RV effect can be sufficiently controlled. In closing, we highlight that the consistency with zero of our measured $b_v$ for the CMASS data has interesting possible implications for galaxy formation models. Our constraint on $b_v$ suggests that galaxies do not have strong memories of their less-massive high redshift progenitors. Given that only a small fraction of the stars in LRGs at $z\sim 0$ were produced in the high-redshift small halos most affected by the relative velocity, this finding is not unexpected. Our constraint suggests that feedback is likely efficient at erasing any differences between galaxies formed in high relative velocity regions and low relative velocity regions. While mergers also play a role in the evolution of small high-redshift halos into the LRGs used for the BAO, the relative velocity's coherence scale is sufficiently large that we do not expect mergers could by themselves erase a relative velocity imprint. Further exploration of this point may be a worthwhile avenue of future work.

1607.06098

1607

1607.08616_arXiv.txt

The circumgalactic medium (CGM) remains one of the least constrained components of galaxies and as such has significant potential for advancing galaxy formation theories. In this work, we vary the extragalactic ultraviolet background for a high-resolution cosmological simulation of a Milky Way-like galaxy and examine the effect on the absorption and emission properties of metals in the CGM. We find that a reduced quasar background brings the column density predictions into better agreement with recent data. Similarly, when the observationally derived physical properties of the gas are compared to the simulation, we find that the simulation gas is always at temperatures approximately 0.5 dex higher. Thus, similar column densities can be produced from fundamentally different gas. However, emission maps can provide complementary information to the line-of-sight column densities to better derive gas properties. From the simulations, we find that the brightest emission is less sensitive to the extragalactic background and that it closely follows the fundamental filamentary structure of the halo. This becomes increasingly true as the galaxy evolves from $z=1$ to $z=0$ and the majority of the gas transitions to a hotter, more diffuse phase. For the brightest ions (C{\scriptsize III}, C{\scriptsize IV}, O{\scriptsize VI}), detectable emission can extend as far as 120 kpc at $z=0$. Finally, resolution is a limiting factor for the conclusions we can draw from emission observations but with moderate resolution and reasonable detection limits, upcoming instrumentation should place constraints on the physical properties of the CGM.

Perhaps the most basic process of galaxy formation, the flow of gas into and out of a galaxy, remains as one of the least understood. The key seems to lie in our lack of understanding of the circumgalactic medium (CGM). Roughly defined as the gas surrounding galaxies at 10 to 300 kpc, the CGM encompasses all gas in transition: gas falling onto the galaxy for the first time; gas that is being driven out by multiple feedback processes; gas that is being stripped from infalling satellite galaxies; and gas that is currently being recycled by the galaxy \citep[see][for review]{Putman_review}. The structure of this gas halo depends on the mass and redshift of the galaxy in question. Currently, gas is thought to be accreted through two main modes - a ``hot'' mode where the gas is shock heated as it enters the halo and a ``cold'' mode where the gas remains in unshocked filamentary structures that can potentially penetrate all the way to the disk \citep{Keres_2005,Fumagalli_2011}. Milky Way-like galaxies are thought to transition from the cold mode to the hot mode by the present day but the details of this transition are neither theoretically agreed upon nor well-constrained observationally \citep{Brooks_2009,ryan,Nelson_2015b}. In addition to these inflows, the outflow of gas from the galaxy is equally important in shaping the CGM \citep{Nelson_2015a, Marasco_2015,Suresh_2015}. Stellar feedback of some form is clearly needed to prevent the overcooling of gas and the formation of unrealistic stellar bulges in simulations \citep{Agertz_2011,Brook_2011,Hummels_2012}. It is also the most effective way of enriching the IGM to the non-pristine levels that are observed \citep{Oppenheimer_2008, Wiersma_2010,Barai_2013,Ford_2013}. While such outflows are regularly seen, the exact physical process driving them and the extent of their influence is uncertain \citep{Turner_2015}. Multiple preferred forms of Type II supernovae feedback are implemented and recent work has begun to implement more detailed processes such as radiation pressure from supernovae \citep{Hopkins_2012,Agertz_2013,Ceverino_2014,Trujillo_Gomez_2015}, cosmic rays \citep{Booth_2013,Salem_2014a,Salem_2014b}, AGN \citep{Sijacki_2007,Booth_2009}, and direct modeling of a kinetic energy component \citep{Simpson_2015} to name a few. In short, putting constraints on these many models is fundamental to furthering our understanding of galaxy formation. In general, cosmological galaxy simulations are tuned to reproduce global and primarily \emph{stellar} properties of galaxies such as the stellar mass function and the star formation rate density function \citep{Dave_2011,schaye_eagle, Nelson_2015a}. Another benchmark is the creation of thin, extended stellar disks \citep{Governato_2007}. The H{\scriptsize I} mass function is a constraining gas property but again looks at the total mass and not its distribution throughout the galaxy. \citep{Dave_2013} Recently, theoretical work has begun to compare the simulated CGM to column densities and equivalent width measurements as a function of impact parameter from the center of the galaxy \citep{hummels, ford, Liang_2015, oppenheimer_2016}. The majority of the simulations have difficulty in matching the large amount and high covering fraction of O{\scriptsize VI} measurements, tracing the hottest gas phase (except recently for high-mass galaxies \citep{Suresh_2015b} and with cosmic ray feedback \citep{Salem_2015}). Their success varies when looking at cooler, less ionized lines (Mg{\scriptsize II}, C{\scriptsize III}, Si{\scriptsize IV} etc.) but in general, the data reveal large amounts of metal-enriched gas at large impact parameters that is hard to reproduce theoretically. In this way, measurements of the CGM can put strong restrictions on feedback models, independent of the global properties that are already used. The most successful method of observing the CGM is in the absorption lines of quasar spectra. At higher redshifts, Lyman $\alpha$ and the ultraviolet metal lines of interest have shifted into the optical, making observations easier and successful \citep{Steidel_2010,Simcoe_2004}. At low-redshift, several studies have begun pushing our knowledge of the more local CGM with measurements of Mg{\scriptsize II} \citep{Chen_2010} and O{\scriptsize VI} for a number of galaxies \citep{Prochaska_2011, Thom_2008}. The recent installation of the Cosmic Origins Spectrograph (COS) on HST has enabled a new survey of the CGM of low-redshift ($z \approx 0.2$), massive, isolated galaxies. The COS Halos Survey has provided a large, uniformly measured sample of the H{\scriptsize I} column densities \citep{Tumlinson_HI}, metal line absorption \citep{werk13}, and O{\scriptsize VI} column densities \citep{tumlinson_OVI}. As accretion and outflows are expected to vary with redshift in addition to mass, low-redshift studies such as these are crucial as is the need to push to even lower redshifts. A complementary approach is to observe the CGM $\emph{directly}$ in emission. Quasar spectra will always be limited by the small number of sight lines through each galaxy. An emission map has the potential to provide insight into the physical state of an entire galaxy halo. While promising, the low density of the gas has made this observation challenging. The most success has come from high-redshift surveys for Lyman $\alpha$ emitters \citep[e.g.][]{Bridge_2013,Gawiser_2007} and the more extended Lyman $\alpha$ blobs/halos \citep[e.g.][]{Matsuda_2011, Steidel_2011, Steidel_2000} but metal-line emission has remained elusive \citep{Battaia_2015}. Recently, the development of new integral field units, MUSE and CWI (and its successor KCWI), now allows for a study of the kinematics of the gas. Early work has already suggested that the absorbers can be linked to global outflows \citep{Swinbank_2015} as well as filamentary inflows \citep{Martin_2014}. At low redshift, the upcoming FIREBall-2 is building upon its predecessor \citep{Milliard_2010} and pushing the boundaries of low surface brightness UV observations. This, in addition to any small or large near-future UV space telescope mean that direct UV observations of the CGM are closer than ever. With these advancements in mind, this work looks to take advantage of new data while preparing for future observations. We take a high-resolution, cosmological, hydrodynamical simulation of a Milky Way-like galaxy and compare it to recent column density data. We then ask what emission we could presume to detect with upcoming facilities. Previous studies of this same simulation provide a solid foundation for this work. \citet{ximena} demonstrated that in-falling satellites provide much of the cold, high metallicity gas found in the halo at $z=0$ whereas \citet{ryan} quantified how much gas of a given temperature is accreted at low-$z$. This existing physical insight allows us to better understand the evolution of the CGM and the contribution of different accretion modes. In this paper, we look to build on this work when interpreting our emission predictions. In Section 2, basics of the simulation used and the photoionization model are summarized. In Section 3, the simulation is compared to column density observations to put empirical constraints on the interpretation of the simulation. In Section 4, the emission signatures of this gas and how they evolve are examined and its observational properties are explored. Finally, the broader context of the work is discussed in Section 5 and the results are summarized in Section 6.

Observing the predicted gas halos of nearby galaxies has long been a goal of observations but the need to study this gas in the space ultraviolet coupled with the diffuse nature of the gas in question has made this challenging. Now, studies of the CGM at low redshift are entering an unprecedented age of sensitivity. Measurements can begin to constrain theoretical prescriptions in simulations as well as discriminate between them. In this work, we vary the EUVB in a high-resolution cosmological simulation of a Milky Way-like galaxy and examine its role in determining how the simulated column densities compare to recent data. We then predict the emission signal expected from such gas for upcoming instrumentation as well as how it varies with redshift and the physical properties of the gas itself. Our main conclusions can be listed as follows: \begin{enumerate} \item Looking at column density maps at $z=0.2$, the largest values for the column densities of all the ions studied here are found in high density filamentary structures. The low-ions (H{\scriptsize I}, C{\scriptsize III}, Si{\scriptsize IV}) are found almost exclusively in these structures while O{\scriptsize VI} is found throughout the halo as its higher ionization energy allows it to exist in the volume-filling hot gas. \item Varying the quasar component of the standard EUVB can significantly change the predicted column densities of the simulation. In particular, lowering this component by a factor of 100 brings the simulation values into much better agreement with the low-ion data of the COS Halos sample. The simulated O{\scriptsize VI} column densities remain too low at all impact parameters compared to the observed values for star-forming galaxies, even with the strongest EUVB. \item Comparing the gas temperature and density in the simulation to that found through \textsc{cloudy} modeling of the COS Halos data shows that the simulation predicts higher temperatures than the data modeling. This demonstrates that it is possible to produce similar column densities from different gas distributions. \item Examining the redshift evolution of the emission reveals that the emission becomes more structured at later times, tracing the remaining high density, low temperature features. This is in contrast to the majority of the gas which shifts to lower densities and higher temperatures from $z=1$ to $z=0$ due to the weakening of cold gas filaments and the progression of supernova-driven winds. \item A surface brightness limit of 100 photons s$^{-1}$ cm$^{-2}$ sr$^{-1}$ should enable a clear detection of emission from the CGM with, C{\scriptsize III} emission extending as far as 100 kpc and O{\scriptsize VI} as far as 50 kpc at $z=0.2$. The predicted extent stays roughly constant for $0.5 < z< 1.0$ as the cosmological surface brightness dimming is balanced by an increasing intrinsic emissivity. \item An angular resolution of 4'' is necessary to begin to resolve the spatial distribution of the CGM out to $z=1$ and sub-arcsecond resolution is needed to resolve beyond a general elongation from the disk. At $z=0.2$, this same observations require an angular resolution of 7.6" (for elongation) and 1.5" (for features) respectively. \end{enumerate} These conclusions focus on results from the combination of predicted UV absorption and emission-line data from a simulated Milky Way-like galaxy, offering a physical explanation for the trends seen in observations for the existence and extent of multiple ions. Other studies have focused on varying feedback prescriptions to bring simulations into better agreement with recent data. However, this can also be reversed as simulation predictions can be extended to create true mock observations that can enable better interpretations of future data. To make more accurate predictions for observations, future work will have to include a number of details excluded here. First, the low surface brightness of the emission in question means that the UV background can overpower the CGM signal. Including a model of the background signal and incorporating its subtraction will provide a better understanding of which CGM structures can be detected with confidence. Second, the continuum emission from the galaxy can also dominate the CGM emission-line signal close to the disk, especially at moderate to low resolution. The disk-halo interface is where the SN winds are being launched; understanding this transition is particularly important. Finally, the velocity structure of the gas has not been considered here, which can change the line profiles of the emission. \citet{ryan} examined the flow of gas into and out of the galaxy, finding that the majority of the accretion at low redshift was in the form of warm/hot gas. Associating emission with these flows is left for future work but will become crucial as integral field units that provide both spatial and spectral information are becoming commonplace. This kinematic information will provide the best observational evidence for both inflows and outflows of gas from galaxies. We are grateful to M. Ryan Joung for generously sharing his simulation output and guidance. L.C. would also like to thank Yuan Li, Cameron Hummels, Munier Salem and Bruno Milliard for helpful discussions. L.C. and D.S. acknowledge support from NASA grant NNX12AF29G. L.C. would also like to acknowledge support from the Chateaubriand Fellowship.

1607.08616

1607

1607.01737_arXiv.txt

A quasi-periodicity has been identified in the strange emission shifts in pulsar B1859+07 and possibly B0919+06. These events, first investigated by Rankin, Rodriguez \& Wright in 2006, originally appeared disordered or random, but further mapping as well as Fourier analysis has revealed that they occur on a fairly regular basis of approximately 150 rotation periods in B1859+07 and perhaps some 700 in B0919+06. The events---which we now refer to as ``swooshes"---are not the result of any known type of mode-changing, but rather we find that they are a uniquely different effect, produced by some mechanism other than any known pulse-modulation phenomenon. Given that we have yet to find another explanation for the swooshes, we have appealed to a last resort for periodicities in astrophysics: orbital dynamics in a binary system. Such putative ``companions'' would then have semi-major axes comparable to the light cylinder radius for both pulsars. However, in order to resist tidal disruption their densities must be at least some 10$^5$ grams/cm$^3$---therefore white-dwarf cores or something even denser might be indicated.

Pulsars are rapidly-rotating, highly-magnetized neutron stars that emit electromagnetic radiation along their magnetic axes which are tilted with respect to their rotation axes. This circumstance causes their radio signals to sweep across our sightline as the star rotates, producing discrete, regular pulses. Such pulse trains or sequences can be complicated by several known phenomena including \textit{nulls} (one or more consecutive pulses with no detectable power) and \textit{mode changes} (abrupt switches between two distinct emission patterns). Both of these effects have been studied extensively and are broadly consistent with our overall understanding of pulsar action--that is, emission forms cones around the magnetic axis that in turn are comprised of a rotating ``beamlet" system. Additional anomalous events were first detected in the emission of radio pulsars B0919+06 and B1859+07 by Rankin, Rodriguez \& Wright (2006, hereafter RRW). In these events, a) the regular pulsed emission gradually moved earlier in longitude over a few pulses, b) remained early for a dozen or more pulses---largely emptying the usual emission window, and c) then returned over a few pulses to the usual longitude. These events attracted notice because they did not seem to be ``classical" mode changes, which exhibit transitions within a single rotation. RRW examined the events in both pulsars, drawing on six full-polarization Arecibo observations (four of B0919+06 and two of B1859+07 made between 1981 and 2005). They attempted to work out the two pulsars' emission geometry and examined several possible explanations for the events. The possibility of some new type of mode changing was examined, but was ruled out because of the gradual onsets and relaxations of the events. Changes in emission altitude were also examined, but the needed displacements within the magnetosphere seemed unreasonably large. Appeals to ``absorption" or other effects seemed ad hoc. In short, RRW were unable to identify a plausible physical cause for these observed events. However, they did show that both the normal and event profiles seemed to stem from partial illumination of the stars' polar caps, as artificial profiles constructed from a comparable number of normal and event pulses produced core-cone triple profiles with the usual core and conal angular dimensions (Rankin 1993, A \& B). Several other studies have drawn attention to these two pulsars. Mitra \& Rankin (2011) conducted sensitive single pulse polarimetry on most of the pulsars originally categorized by Lyne \& Manchester (1988) as ``partial cones"---including B0919+06---and their analysis tends to support the interpretation of RRW to the effect that the events entail a partial illumination of the pulsar's emission cone, late for the normal emission and earlier during the events. Evidence for long term periodicities on the order of tens of years in the timing of B0919+06 were developed by Lyne \etal\ (2010) and Shabanova (2010) and thoroughly studied recently by Perera \etal\ (2015). A similar study on B1859+07, again drawing on the unexcelled pulsar timing archive at the Jodrell Bank Observatory, has been carried out by Perera \etal\ (2016). Most recently, Han \etal\ (2016) published the results of $\sim30$ hours of B0919+06 observations using the Jiamusi 66-m Telescope and found events similar to those identified by RRW. In this paper we undertake a new study of these emission events in pulsars B0919+06 and B1859+07. In order to emphasize their distinct character, we refer to the events as ``swooshes," returning to the (unpublished) terminology used by RRW in the preparation of their paper. First noticed in the 1990s in Arecibo observations of B0919+06, the phenomenon was ignored until it was again seen by RRW---and now in this work with three different Arecibo instruments. That the events were also seen with the Jiamusi Telescope rules out any instrumental cause. After initial analysis of these two pulsars by RRW, it had become clear that an entirely new approach was needed to investigate and explain the swoosh phenomenon, and we began by carrying out much longer and more extensive Arecibo observations of both pulsars. Our analyses produced evidence of a perplexing periodicity in both stars, which in turn opened up new potential avenues of interpretation. \S 2 describes the observations. \S 3 and \S 4 examine events in each star in more detail, and \S 5 and \S 6 discuss the evidence for their periodicity. \S 7, \S 8, and \S 9 then discuss possible mechanisms for the swooshes. Finally \S 10 summarizes the results and discussion. \begin{figure} \begin{center} \hbox{\hspace{-0.55 cm}\includegraphics[width=90mm, angle=0.]{Figure2.pdf}} \vspace{-1.1 cm} \caption{Higher resolution 200-pulse display showing the second B0919+06 swoosh in Figure~\ref{fig1}.} \label{fig2} \end{center} \end{figure} \begin{table} \begin{center} \caption{Observational parameters} \begin{tabular}{cccccc} \hline \hline Pulsar & Date & MJD & Length & Number \\ RF (MHz) & (m/d/yr) & & (pulses) & of Events \\ \hline {\ \bf B0919+06} \\ 1425 & 8/03/2003 & 52854 & 1115 & 2\\ 327 & 10/04/2003 & 52916 & 4180 & 1\\ 325.85 & 5/05/2013 & 56417 & 17259 & 6\\ 325.85 & 10/11/2013 & 56576 & 17810 & 7\\ 1392.5 & 10/12/2013 & 56577 & 13994 & 4\\ 1494 & 8/23/2015 & 57257 & 16558 & 6\\ \\ 2250 & 4/18/2015 & 57130 & 2088 & 1\\ 2250 & 7/12/2015 & 57215 & 53053 & 22\\ 2250 & 7/14/2015 & 57217 & 67611 & 21\\ 2250 & 8/17/2015 & 57251 & 65003 & 26\\ 2250 & 8/18/2015 & 57252 & 35306 & 16\\ 2250 & 8/19/2015 & 57253 & 25001 & 6\\ \\ {\bf B1859+07} \\ 1525 & 04/10/2003 & 52739 & 1021 & $\sim$7\\ 1520 & 1/2/2005 & 53372 & 2096 & $\sim$12\\ 1392.5 & 3/26/2013 & 56377 & 6460 & $\sim$40\\ 1392.5 & 4/8/2015 & 57121 & 12195 & $\sim$171\\ 1394 & 12/5/2015 & 57361 & 12154 & $\sim$141\\ \hline \label{Table1} \end{tabular} \end{center} \normalsize \end{table}

We report extensive new Arecibo observations and analyses studying the strange swooshes in pulsars B0919+06 and B1859+07. We find much more complexity that was earlier reported by Rankin \etal\ (2006); however, we confirm the basic effect involving usually gradual, temporary, about 10\degr\ shifts in the emission window to earlier longitudes. Here also, we provide some evidence for a quasi-periodicity in pulsar B1859+07 and perhaps in B0919+06. We examine a number of effects that might produce the swooshes and find no obvious mechanism. In particular, the effects might represent mode changes, but no example of a gradual mode change is known, so swooshes appear to be some other phenomenon. These difficulties have led us to explore whether some type of co-orbiting companion with an unprecedentedly short period and separation might be involved. Such a system would entail so many effects of tide and heating that it is difficult to predict what the consequences might be. Given that WD-NS and WD-WD binaries are known with periods 10-100 times larger one can ask what happens as those systems evolve.

1607.01737

1607

1607.06117_arXiv.txt

We have conducted a near-infrared (NIR) proper motion survey, the Sondage Infrarouge de Mouvement Propre, in order to discover field ultracool dwarfs (UCD) in the solar neighborhood. The survey was conducted by imaging $\sim28\%$ of the sky with the Cam\'era PAnoramique Proche-InfraRouge both in the southern hemisphere at the Cerro Tololo Inter-American Observatory 1.5 m telescope, and in the northern hemisphere at the Observatoire du Mont-M\'egantic 1.6 m telescope and comparing the source positions from these observations with the Two Micron All-Sky Survey Point Source Catalog (2MASS PSC). Additional color criteria were used to further discriminate unwanted astrophysical sources. We present the results of an NIR spectroscopic follow-up of 169 M, L, and T dwarfs. Among the sources discovered are 2 young field brown dwarfs, 6 unusually red M and L dwarfs, 25 unusually blue M and L dwarfs, 2 candidate unresolved L+T binaries, and 24 peculiar UCDs. Additionally, we add 9 L/T transition dwarfs (L6--T4.5) to the already known objects.

Ever since the discovery of the high proper motion of Barnard's star (\citealt{bar16}), proper motion has proved to be an effective method to identify faint stars in the solar neighborhood. Even thin disk stars, which are comoving with the Sun, have residual velocities of the order of tens of km\,s$^{-1}$. This arises mainly from the fact that, as stars age, they randomly interact with interstellar cloud complexes and spiral arms, which progressively increases their velocity dispersion. This leads to proper motions of about 100\,mas\,yr$^{-1}$ for stars up to $\sim$50\,pc. With the advent of sensitive near-infrared (NIR) hybrid mosaic detectors, the search for low-mass stars and substellar objects has become possible. Hundreds of ultracool dwarfs (UCDs; spectral type M7 or later; \citealt{kir95}) have been discovered through large-scale red-optical (0.6--1.0$\,\mu$m) surveys, such as the Sloan Digital Sky Survey (SDSS; \citealt{yor00}), the Canada--France Brown Dwarf Survey (\citealt{del08}), the Pan-STARRS1 3$\pi$ Survey (PS1; \citealt{kai10}), along with large-scale and all-sky NIR (1.0--2.5$\,\mu$m) surveys, e.g.,~the DEep Near Infrared Survey (DENIS; \citealt{epc97}), the Two Micron All Sky Survey (2MASS; \citealt{skr06}), the UKIRT Infrared Deep Sky Survey (UKIDSS; \citealt{law07}), the VISTA Hemisphere Survey (VHS; \citealt{mcm12}) and the all-sky mid-infrared (3.4--22$\,\mu$m) {\it Wide-Field Infrared Survey Explorer} ({\it WISE}; \citealt{wri10}) survey. Most of these surveys used color-based criteria to identify UCDs (\citealt{del97}; \citealt{kir99}; \citealt{bur02}; \citealt{haw02}; \citealt{kir11}). While this is an effective way of finding UCDs, it may overlook objects with peculiar colors, such as very low metallicity halo objects (subdwarfs; \citealt{bur03}; \citealt{siv09}). Furthermore, an increase in opacity of the CH$_4$, H$_2$O, and H$_2$ molecules in the atmosphere of L/T transition dwarfs (L6--T4.5) leads to NIR colors similar to those of M and early-L dwarfs (\citealt{leg00}; \citealt{chi06}). The ability to select transition dwarf candidates for spectroscopic follow-up is important to uncover a large sample of these objects and better understand the processes causing condensate cloud disruption (\citealt{ack01}; \citealt{bur02b}; \citealt{kna04}), which is known to happen in their photosphere as they evolve and cool down. This effect has been demonstrated in several variability studies (\citealt{cro14} and references therein). While recent studies were able to discover some L/T transition dwarfs (\citealt{bes13}; \citealt{sch14}), it is likely that several more are still hiding in surveys such as 2MASS and {\it WISE}. This could also be said of UCDs very close to the Sun. Previous studies (\citealt{met08}; \citealt{burn10}; \citealt{rey10}; \citealt{kir12}) have suggested that the space density of brown dwarfs is below that of stars, with possibly a minimum at the L/T transition (\citealt{day13} and references therein). However, the recent finding of the brown dwarf binary WISE~J104915.57$\relbar$531906.1 (\citealt{luh13}), and of the coldest brown dwarf known to date, WISE~J085510.83$\relbar$071442.5 (\citealt{luh14}), outclassing Wolf 359 and Lalande 21185 as the third and fourth closest systems to the Sun, provides evidence that the search for UCDs in the solar neighborhood is far from over. The proper motion method is the most efficient way of uncovering nearby ($\lesssim$ 50\,pc) UCDs in a color-unbiased search, just as it is for warmer stars. While some surveys include multi-epoch observations that can lead to a detection of proper motion (\citealt{loo08}; \citealt{kir10}), the small timespan (approximately one to two years) between individual observations is better suited for the detection of very large proper motions. This is the case for AllWISE (\citealt{kir14}), released to track down high-proper motion objects in the {\it WISE} catalog. However, as noted by \citet{kir14}, AllWISE measures only the apparent motion on the sky, because most objects were observed only at two epochs separated by six months. This interval maximizes the effect of the parallax on the motion of the sources, hindering the calculation of their true proper motion. Some studies have combined the different epochs of two or more of those surveys to search for high proper motion objects: DENIS and 2MASS (\citealt{art10}), 2MASS and SDSS (\citealt{met08}; \citealt{she09}; \citealt{gei11}), 2MASS and UKIDSS (\citealt{dea09}), and more recently 2MASS and {\it WISE} (\citealt{bih13}; \citealt{per14}) or 2MASS and AllWISE (\citealt{gag15}) and Pan-Starrs and AllWISE (\citealt{bes13}). Others used several of those surveys to identify new candidates (\citealt{abe11}; \citealt{sch12}) or added a second epoch of $K$-band astrometry, such as the UKIDSS Galactic Plane Survey (GPS; \citealt{luc08}), in which a new T5 dwarf and several other L and T dwarf candidates were recently discovered in the galactic plane (\citealt{smi14}). It is in this context that the large area NIR proper motion survey Sondage Infrarouge de Mouvement Propre (SIMP) was initiated using the 2MASS Point Source Catalog (PSC\footnote{see \protect\url{http://irsa.ipac.caltech.edu/}}) as a first epoch. This brings the possibility of discovering UCDs in the solar neighborhood that were missed in other searches, such as L/T transition dwarfs or UCDs with peculiar photometric colors. This survey is also suited for the discovery of wide binaries and members of young moving groups. Discovering more of these objects will help constrain the substellar mass function and the brown dwarf space density in the solar neighborhood. Furthermore, cataloging a large number of high-proper motion brown dwarfs will be useful for predicting microlensing events, enabling an independent and direct measurement of their masses (\citealt{eva14}). Since 2MASS and SIMP use similar filters and pixel scales and reach photometric comparable depths (see Section~\ref{SIMPsurvey}), cross-matching sources is straightforward. This survey has already enabled the discovery of several UCDs: a bright, nearby T2.5 dwarf (SIMP~J013656.5+093347; \citealt{art06}), five very low-mass binaries (SIMP~J1619275+031350AB $\&$ SIMP~J1501530$\relbar$013506AB; \citealt{art11}, 2MASS~J10432513$\relbar$1706065, 2MASS~J11150150+1607026 $\&$ 2MASS~J12594167+1001380; \citealt{bar15}), and a low-gravity L5~$\beta/\gamma$ candidate member of the Argus association (SIMP~J21543454$\relbar$1055308; \citealt{gag14, gag15b}). This paper presents the remaining UCDs discovered so far through the SIMP survey.\footnote{A list of all known ultracool dwarfs is maintained at \protect\url{http://www.astro.umontreal.ca/~gagne/listLTYs.php}.} The SIMP survey is described in detail in Section~\ref{SIMPsurvey}, and the spectroscopic and photometric follow-ups are presented in Section~\ref{spectrocand} and Section~\ref{PhotoObs}, respectively. We describe the method used to classify the candidates in Section~\ref{spectral_class} and we present the results in Section~\ref{Results}. We summarize and conclude in Section~\ref{Summary}.

\label{Summary} A sample of 169 spectra have been obtained from different observatories for UCD candidates found through proper motion detection by comparing source position from the SIMP survey to 2MASS. From this sample, 164 UCDs, with spectral types from M8 to T6.5, were found, including 2 young field brown dwarfs, 6 unusually red M and L dwarfs, 25 unusually blue M and L dwarfs, 2 candidate unresolved L+T binaries and 24 other peculiar UCDs. We also discovered 9 L/T transition dwarfs (L6-T4.5), which helps better populate this interesting range of spectral type, often the target of variability studies. This work demonstrates that many undiscovered UCDs remain to be found in preexisting surveys, including several peculiar objects. Proper motion has proved to be an efficient method to uncover those missed objects. Furthermore, most objects found are relatively bright ($J < 16$), enabling follow-up studies (parallax, variability, binarity, etc.).

1607.06117

1607

1607.05925_arXiv.txt

{HBC\,722 (V2493\,Cyg) is a young eruptive star in outburst since 2010. Spectroscopic evidences suggest that the source is an FU Orionis-type object, with an atypically low outburst luminosity.} {Because it was well characterized in the pre-outburst phase, HBC\,722 is one of the few FUors where we can learn about the physical changes and processes associated with the eruption, including the role of the circumstellar environment.} {We monitored the source in the $BVRIJHK_S$ bands from the ground, and at 3.6 and 4.5\,$\mu$m from space with the Spitzer Space Telescope. We analyzed the light curves and studied how the spectral energy distribution evolved by fitting a series of steady accretion disk models at many epochs covering the outburst. We also analyzed the spectral properties of the source based on our new optical and infrared spectra, comparing our line inventory with those published in the literature for other epochs. We also mapped HBC\,722 and its surroundings at millimeter wavelengths.} {From the light curve analysis we concluded that the first peak of the outburst in 2010 September was mainly due to an abrupt increase of the accretion rate in the innermost part of the system. This was followed after a few months by a long term process, when the brightening of the source was mainly due to a gradual increase of the accretion rate and the emitting area. Our new observations show that the source is currently in a constant ``plateau'' phase. We found that the optical spectrum was similar both in the first peak and the following periods, but around the peak the continuum was bluer and the H$\alpha$ profile changed significantly between 2012 and 2013. The source was not detected in the millimeter continuum, but we discovered a flattened molecular gas structure with a diameter of 1700\,au and mass of 0.3\,M$_{\odot}$ centered on HBC~722.} {While the first brightness peak could be interpreted as a rapid fall of piled-up material from the inner disk onto the star, the later monotonic flux rise suggests the outward expansion of a hot component according to the theory of Bell \& Lin (1994). Our study of HBC\,722 demonstrated that accretion-related outbursts can occur in young stellar objects even with very low mass disks, in the late Class\,II phase.}

Sun-like pre-main sequence stars are surrounded by circumstellar disks, from which material is accreted onto the growing protostar. The accretion rate is variable: the protostar's normal accretion at a low rate may be occasionally interspersed by brief episodes of highly enhanced accretion \citep{kenyon1990}. FU Orionis-type variables (FUors) are thought to be the visible examples of episodic accretion. During their episodic ``outbursts'', accretion rate from the circumstellar disk onto the star increases by several orders of magnitude, from typically $10^{-7}$ up to 10$^{-4}$\,M$_{\odot}$\,yr$^{-1}$ \citep{audard2014}. Due to the increased accretion, FUors brighten by 5--6\,mag at optical wavelengths, their bolometric luminosities reach several hundred L$_{\odot}$, and they stay in the high state for several decades. Currently, about two dozen FUors and FUor candidates are known. One of the recent discoveries, HBC\,722 went into eruption in 2010 \citep{semkov2010a}. The object's optical and near-infrared (near-IR) spectra published by \citet{miller2011} and by \citet{semkov2012} are similar to those of FU\,Ori and other FUors, confirming its FUor-type classification. HBC\,722, however, differs from typical FUors in that its outburst luminosity and accretion rate are only on the order of $L_{\rm bol}$ = 10--20\,L$_{\odot}$, $\dot{M}$ = 10$^{-6}$\,M$_{\odot}$\,yr$^{-1}$, well below what is usual for FUors \citep[][hereafter, Paper\,I]{kospal2011}. Despite this, HBC\,722 has been in the bright state for at least five years now, it is currently brighter than ever \citep{baek2015}, and it is on its way to exhibit a typical, decades-long FUor-type light curve. HBC\,722 is not an isolated object but a part of the LkH$\alpha$ 188 cluster, a group of optically visible young stars showing H$\alpha$ emission \citep{cohen1979}. Several Class 0/I embedded protostars, as well as very young, very low luminosity protostars or starless cores in the vicinity of HBC\,722 indicate active star formation in the area \citep{green2011,dunham2012b}. HBC\,722 itself is a Class\,II object with a circumstellar disk \citep{miller2011,kospal2011}, although with a rather low upper limit for the disk mass \citep{dunham2012b}. The system is surrounded by a reflection nebula \citep{miller2011}. HBC\,722 is unique among FUors in that the progenitor, i.e., the object in quiescence, has been well characterized. In order to learn about the physical changes and processes associated to the outburst, we carried out new optical and infrared photometric monitoring of HBC\,722, including mid-IR observations, which provide new information on the thermal emission of the inner disk at various phases of the outburst. We also obtained optical and near-IR spectra, as well as millimeter continuum and molecular line maps of the environment of the source, and compared our results with the FUor outburst theory of \citet{bell1994}.

\label{sec:con} Based on our optical-infrared monitoring, spectroscopic observations, and millimeter mapping, we performed a detailed study of the first six years of the outburst of the low-mass T\,Tauri star HBC\,722. We interpreted the first brightness peak, lasting several months, as the rapid fall of piled-up material from the inner disk onto the star. This was followed by a monotonic flux rise, which can be explained by increasing accretion rate and emitting area. Our observations are consistent with the predictions of \citet{bell1995} for a system where thermal instability is triggered at an intermediate radius within the metastable disk area. Their model predicts a rapid ionization from propagating toward the star, causing a short-lived brightness peak, similar to what we observed in 2010 September. In parallel, a second, slower ionization front starts to expand outward, which can be the physical reason behind the re-brightening of HBC~722 after 2011 September. Our study of HBC\,722 demonstrated that accretion-related outbursts can occur in young stellar objects even with very low mass disks. Our results strengthen the theory that eruptive phenomena may appear throughout star formation from the embedded phase to the Class\,II phase, and that possibly all young stars undergo phases of temporarily increased accretion.

1607.05925

1607

1607.02996_arXiv.txt

{Solar prominences are subject to both field-aligned (longitudinal) and transverse oscillatory motions, as evidenced by an increasing number of observations. Large-amplitude longitudinal motions provide valuable information on the geometry of the filament-channel magnetic structure that supports the cool prominence plasma against gravity. Our pendulum model, in which the restoring force is the gravity projected along the dipped field lines of the magnetic structure, best explains these oscillations. However, several factors can influence the longitudinal oscillations, potentially invalidating the pendulum model. }{ The aim of this work is to study the influence of large-scale variations in the magnetic field strength along the field lines, i.e., variations of the cross-sectional area along the flux tubes supporting prominence threads. }{ We studied the normal modes of several flux tube configurations, using linear perturbation analysis, to assess the influence of different geometrical parameters on the oscillation properties. }{ We found that the influence of the symmetric and asymmetric expansion factors on longitudinal oscillations is small.}{We conclude that the longitudinal oscillations are not significantly influenced by variations of the cross-section of the flux tubes, validating the pendulum model in this context.}

\label{sec:intro} Solar prominences are subject to various types of oscillatory motions \citep[see review by][]{arregui2012}. In general these oscillations are a combination of movements transverse and longitudinal to the local magnetic field. In early magnetohydrodynamic (MHD) descriptions of longitudinal oscillations, the restoring force was proposed to be gas pressure gradients, as in the so-called slow modes. However, our numerical simulations have shown that large-amplitude longitudinal oscillations (LALOs) are strongly influenced by the solar gravity projected along the dipped field lines that support the prominence \citep{luna2012b}. In fact, gravity is the most important restoring force under prominence conditions, according to our pendulum model, and the oscillation period depends mainly on the curvature of the field lines. Subsequent parametric numerical studies of longitudinal oscillations for different tube geometries found good agreement with our pendulum model \citep{zhang2012,Zhang2013a}. In \citet{luna2012c}, hereafter called Paper I, we studied the effects of the field-line curvature on the longitudinal linear modes in flux tubes with uniform cross-sections (i.e., uniform magnetic field) but with different geometries. We found that slow modes become magnetoacoustic-gravity modes when dipped field lines are considered instead of horizontal straight lines. In those modes the restoring force is a combination of gravity projected along the field lines and gas pressure gradients, however, under typical prominence conditions the gravity dominates and is the main restoring force. Here we extend the analysis of Paper I by including the variations of the magnetic field strength along the prominence-supporting flux tubes and determining their influence on the magnetosonic-gravity modes. These variations are converted to cross-sectional area variations according to conservation of magnetic flux: $B(s) \, A(s) = \mathrm{constant}$, where $B(s)$ and $A(s)$ are the magnetic field and cross-sectional area along the tube, and $s$ is the curved coordinate along the tube. In the governing equations \citep[see, e.g.,][]{goossens1985,karpen2005}, derivatives of the magnetic field along the tube appear explicitly. Therefore longitudinal oscillations can be influenced by variations of the flux tube cross-section. This study establishes the effects of the flux tube cross-section on LALOs and the circumstances under which these effects are significant. In addition, this work contributes to prominence seismology because the results can be used to determine the magnetic field variations along the dipped field lines in observed prominence threads undergoing LALOs. There are two main drawbacks to our model for the LALOs. We consider a flux tube as an isolated entity, and assume that the longitudinal oscillations are completely decoupled from the transverse oscillations. In reality, however, the flux tube is embedded in a large filament-channel structure, so the oscillations might depend on the global structure instead of the properties of isolated field lines. In addition the longitudinal and transverse motions might be coupled. However, studies of the linear global oscillations of a two-dimensional (2D) prominence structure revealed that, for the relatively small plasma-$\beta$ (ratio of gas pressure to magnetic pressure) typical for prominences, slow modes are decoupled from fast modes and are associated mainly with longitudinal motions \citep{joarder1992,oliver1993}. \citet{goossens1985} studied oscillations of a 2.5D magnetic structure in a general situation of an arbitrary plasma-$\beta$. The authors determined the oscillatory spectrum of longitudinal, fast, and Alfv\'en modes. Part of both the longitudinal and Alfv\'en spectrum are formed of continuous modes. These continuous spectra are given by an eigenvalue problem on each magnetic surface and depend on the equilibrium quantities in different magnetic surfaces. This means that two different magnetic surfaces oscillate with different frequencies. In contrast, the discrete spectrum consists of modes of the whole structure where all the points oscillate with the same frequency. In general, both continua of Alfv\'en and longitudinal modes are coupled. However, the authors found that the two continua are not coupled if the magnetic structure is purely poloidal, i.e., the magnetic field is contained in the $xz$-plane and $B_y=0$. This is the situation considered in this work and justifies our treatment of the longitudinal oscillations in flux tubes as isolated structures. In addition, we are not interested in global modes or in the motions along the ignorable direction related to the Alfv\'en oscillations, rather we focus on the longitudinal continuum modes. More recently, we studied nonlinear longitudinal and transverse motions in a 2D prominence structure \citep{luna2016}. We found that the longitudinal and transverse oscillations are not coupled, even in the nonlinear regime. The longitudinal motions agree very well with our pendulum model \citep{luna2012b}. In particular the plasma contained in each dipped flux tube oscillates with its own longitudinal frequency that depends mainly on the curvature of the field line dip, forming a continuum of frequencies. We also found that the transverse motions of the different field lines are identical, indicating that a global fast normal mode is established. This result confirms the validity of studying LALOs in prominences with the isolated flux tube of Paper I and the present work. In Section \ref{sec:model} the model is presented, governing equations are described, and normal modes of the system are derived analytically for different flux-tube geometries. In Section \ref{sec:sym_tubes} the solutions for symmetric flux-tubes are shown. In Section \ref{sec:anti_tubes} the normal modes of asymmetric tubes are presented, whereas in Section \ref{sec:general_case} the mixed tubes are studied. Finally, in Section \ref{sec:conclusions} we summarize the results of this investigation and our conclusions.

\label{sec:conclusions} In this work we investigated the influence of the cross-sectional area, i.e., magnetic field, variations on the frequencies and velocity profiles of the normal modes in a flux tube containing a prominence thread surrounded by hot corona. We characterized the area variations with two parameters: the symmetric and asymmetric terms that we called expansion parameters. The asymmetric term produces flux tubes with cross-sections that are not symmetric about the tube midpoint, i.e., with one end of the tube that is thicker than the other. The symmetric term produces tubes with symmetric cross-sections, where both ends have the same area; in this case the midpoint of the tube is thinner than the ends for positive values and thicker for negative values of the symmetric expansion factor. Our model prominence flux tube is segmented into three isothermal regions: a cool thread with a symmetric or asymmetric area expansion in the center, bounded by two horizontal side segments of hot coronal plasma with no area expansion. We derived the equation that describes the linear perturbations of the velocity along the model flux tube, and solved for the normal modes for symmetric, asymmetric, and mixed flux-tube geometries. In general, the normal modes lack symmetry with respect to the tube midpoint and can be classified into two categories. The first-kind modes have an even number of nodes (zeroes) of the velocity along the entire tube, whereas the second-kind have an odd number of nodes. The fundamental mode, which corresponds to the pendulum mode of Paper I, is a first-kind mode in all cases considered. We found that the expansion factors generally exert little influence on the longitudinal oscillation frequencies. For small radius of curvature and large temperature contrast, the influence of the cross-sectional variation increases although not significantly. In a tube described by a combination of the symmetric and asymmetric expansion factors the influence is in some cases slightly larger. For example, in the limit of almost straight field lines (large $R$), the frequencies diverge by $<9\%$ from the uniform-tube frequencies. The shape of the flux tube also affects the spatial profile of the normal-mode velocity. For the fundamental mode, the symmetric expansion factor produces a relatively small enhancement of the velocity in the center of the tube. In contrast, the asymmetric expansion term clearly produces different velocities at the thread ends by as much as a factor of 3 for realistic expansion factors. This is the most significant result in terms of observable signatures that could potentially provide insight into the geometry of prominence-bearing flux tubes. We conclude that, in general, the longitudinal oscillations are not strongly influenced by nonuniform magnetic field strength, i.e., cross-sectional area, along the flux tube. Relatively high values of the asymmetric expansion term, however, can produce significant velocity differences at both ends of the prominence threads. These results confirm that the pendulum model of Paper I remains valid for flux tubes with a range of cross-sectional area variations that are consistent with observations, but the model considered here does not allow interactions between the plasma and magnetic field. Recently we simulated the effects of perturbing localized cool plasma supported by a 2D dipped magnetic field \citep{luna2016}. We found that the back-reaction of the field to the plasma oscillation is very small, validating the simpler assumption of rigid flux tubes in Paper I and the present study. Nonlinear 3D simulations \citep[e.g.,][]{xia2016} are needed to fully understand the oscillatory modes of prominence plasma embedded in the hot, magnetized corona, particularly for the large-amplitude longitudinal oscillations commonly observed near energetic events \citep[e.g.,][]{luna2014}.

1607.02996

1607

1607.06466_arXiv.txt

We obtained adaptive-optics assisted \S\ observations of the central regions of the giant elliptical galaxy \n5419 with a spatial resolution of $0.2$~arcsec ($\approx 55$~pc). \n5419 has a large depleted stellar core with a radius of 1.58~arcsec (430~pc). \HST\ and \S\ images show a point source located at the galaxy's photocentre, which is likely associated with the low-luminosity AGN previously detected in \n5419. Both the \HST\ and \S\ images also show a second nucleus, off-centred by 0.25~arcsec ($\approx 70$~pc). Outside of the central double nucleus, we measure an almost constant velocity dispersion of $\sigma \sim 350$\kms. In the region where the double nucleus is located, the dispersion rises steeply to a peak value of $\sim 420$\kms. In addition to the \S\ data, we also obtained stellar kinematics at larger radii from the South African Large Telescope. While \n5419 shows low rotation ($v < 50$\kms), the central regions (inside $\sim 4 \, r_b$) clearly rotate in the opposite direction to the galaxy's outer parts. We use orbit-based dynamical models to measure the black hole mass of \n5419 from the kinematical data outside of the double nuclear structure. The models imply $\mbh=7.2^{+2.7}_{-1.9} \times 10^9$~\Ms. The enhanced velocity dispersion in the region of the double nucleus suggests that \n5419 possibly hosts two supermassive black holes at its centre, separated by only $\approx 70$~pc. Yet our measured $\mbh$ is consistent with the black hole mass expected from the size of the galaxy's depleted stellar core. This suggests, that systematic uncertainties in $\mbh$ related to the secondary nucleus are small.

It is now well established that super massive black holes (SMBHs) are present at the centres of galaxies with bulges. The observational correlations found between the black-hole mass and various parameters of the host galaxy, e.g., velocity dispersion, bulge mass, and bulge luminosity, have led to the idea that the formation and evolution of early type galaxies and their nuclear SMBHs are tightly linked \citep[see e.g.][and references therein]{Kormendy2013}. \n5419 is a luminous elliptical galaxy ($M_V = -23.1$). Such bright early type galaxies (brighter than $M_V \sim -21$) often have a low density `core' with a typical size of a few tens or hundreds of parsecs. Inside the core or break radius, $r_b$, the light profile is much shallower than the inward extrapolation of the outer S\'ersic profile \citep[e.g.][]{Lauer1995,Lauer2005,Graham2003a,Rusli2013b}. Moreover, core galaxies differ from fainter ellipticals in their isophotal shapes and degree of rotational support: core ellipticals have boxy instead of disky isophotes and are supported by anisotropic velocity dispersions rather than rotational stellar motions \citep[e.g.][]{Nieto1991a,Kormendy1996a,Faber1997,Lauer2012a}. These morphological and kinematic distinctions between core galaxies and fainter ellipticals suggest that the processes involved in their formation are different. The most plausible mechanism for core formation is black hole scouring that occurs in non-dissipative mergers of galaxies as a result of the dynamics of their central SMBHs: the SMBHs spiral into the centre of the merger via dynamical friction, ultimately forming a black hole binary \citep[][]{Begelman1980}. As the binary shrinks, it ejects stars on intersecting orbits, creating a low-density core with a mass deficit of the order of or a few times larger than the mass of the binary \citep[see e.g.][and references therein]{Milosavljevic2001a,Merritt2006a,Merritt2013a}. The black hole binary model can explain the correlations between core structure, black hole masses, mass deficits and the observed orbital structure in core galaxies \citep[e.g.][]{Faber1997,Milosavljevic2001a, Kormendy2009a, Hopkins2009a,Thomas2014,Thomas2016}. Dissipation-driven core formation has been reported from numerical $N$-body simulations that include the dynamical effects of AGN feedback, but these results have not been tested yet in detail against the wealth of observations available for core elliptical galaxies. Core scouring implies the formation of black hole binaries. Their subsequent evolution can follow different paths. In particular, if the binary separation decreases enough and the SMBHs manage to coalesce, the merged black hole could be ejected from the galaxy centre by anisotropic emission of gravitational waves \citep[e.g.][]{Begelman1980}. As the probability that the remnant black hole recoils at a velocity exceeding the escape velocity of a large elliptical galaxy is low \citep[e.g.][]{Lousto2012a}, it will most likely remain bound to the galaxy on a radial orbit. On the other hand, if the binary decay stalls, subsequent mergers may bring a third SMBH or second black hole binary to the centre \citep[e.g.][]{Valtonen1996}. The interactions of this newly formed multiple SMBH system will eventually displace one or more of the SMBHs from the nucleus. In both of these scenarios, off-centred SMBHs are expected to be found in the centre of bright elliptical galaxies. The observational evidence of binary or recoiled SMBHs that supports these theoretical predictions is still scarce, usually involving pairs of SMBHs at kiloparsec separations \citep[e.g.][]{Komossa2003a,Ballo2004,Bianchi2008,Koss2011,Fu2011a,McGurk2015,Comerford2015}. On smaller scales, mostly indirect evidence has been reported, with only one secure case known so far \citep[CSO 0402$+$379, ][]{Rodriguez2006}. In this paper we study \n5419, the dominant galaxy of the poor cluster Abell~753. This large elliptical galaxy was first classified as a core by \citet{Lauer2005}, who modelled its inner $\sim 10$~arcsec surface brightness profile derived from a {\it HST}/WFPC2 F555W image. Based on a Nuker-law \citep{Lauer1995} fit to \n5419 profile, they derived a break radius $r_b = 2.38$~arcsec (650~pc), making this relatively low-surface brightness core the largest among the 42 objects classified as core galaxies in their sample. Moreover, the {\it HST} image presented by \citet[][]{Lauer2005} revealed the presence of a double nucleus at the centre of the galaxy, with a projected separation of a few tens of parsecs (see also \citealt{Capetti2005b,Lena2014}; this work). Additionally, radio observations \citep[][]{Goss1987,Subrahmanyan2003} and the detection of hard X-rays \citep[][]{Balmaverde2006} indicate that the centre of \n5419 hosts a low-luminosity AGN (LLAGN). All this makes \n5419 an interesting case for studying the interplay between core galaxies and their SMBHs. We observed \n5419 as part of our black hole survey, consisting of 30 galaxies observed with \S\ at the Very Large Telescope (VLT, \citealt{Nowak2007, Nowak2008, Nowak2010, Rusli2011, Rusli2013a, Rusli2013b, Saglia2016}; Bender et al. in preparation; Erwin et al. in preparation; Thomas et al. in preparation). Here we report the results obtained for \n5419. The \S\ data of \n5419 and additional long-slit spectroscopy from the Southern African Large Telescope are presented in Section~\ref{s_data}. In Sections~\ref{s_photo} and \ref{s_kin} we present the results from the imaging and spectroscopy of \n5419. Section~\ref{s_dyn} deals with the dynamical modelling of the galaxy. A discussion about the double nucleus of \n5419 can be found in Section~\ref{s_nucleus} and our final conclusions in Section~\ref{s_summary}. We assume a distance to \n5419 of 56.2~Mpc, derived from the radial velocity corrected for Local Group infall onto Virgo taken from Hyperleda, $v_{\rm vir}=4047$\kms, and a value of $H_0 = 72~{\rm km~s^{-1}~Mpc^{-1}}$. At this distance, 1~arcsec corresponds to 273~pc.

\label{s_summary} We have presented high-resolution \Kb\ \S/VLT IFS and optical long-slit spectroscopic observations of the galaxy \n5419. The kinematics derived from these data show that \n5419 is a dispersion-dominated galaxy. The rotational velocity does not exceed 50\kms; although a clear rotational pattern is observed, revealing a counter-rotating core in the inner few arcseconds. The velocity dispersion is about 350\kms and almost constant over the low-surface brightness core. However, inside $0.35$~arcsec ($\approx 100$~pc), where the galaxy hosts a double nucleus, it increases reaching values of 420-430\kms. We use orbit-based dynamical models to model the stellar kinematics outside the double nucleus in the centre. From this analysis we derive a \Mbh\ of $7.24^{+2.74}_{-1.91} \times 10^9$~\Ms. This mass is consistent with the large core radius ($r_b=1.58$~arcsec or $\approx 430$~pc) obtained by fitting a Core-S\'ersic function to the surface brightness profile, given the known correlation between core radius and SMBH mass \citep{Rusli2013b}. The \Rb\ mass-to-light ratio derived from the dynamical modelling, $\ml = 5.37$, is about a factor $1.1-1.4$ larger than the one derived from the stellar population analysis in the inner few arcseconds, for which we assumed a Kroupa IMF. The dynamical $\ml$ thus lies between a Kroupa and Salpeter IMF. We have also discussed different scenarios in which the observed properties of the double nucleus of \n5419 can be explained. While the nature of the double nucleus in \n5419 is certainly puzzling, our observations suggest that this galaxy might host two SMBHs in close proximity. If this is the case, \n5419 is a promising target to study the interaction between SMBHs at the centre of galaxies, and their formation and evolution. More clues could be obtained from milliarcsecond radio imaging and higher spatial-resolution spectroscopic data, that would allow us to study in a resolved manner the structures seen in the central 0.5~arcsec (135~pc) of \n5419.

1607.06466

1607

1607.03133_arXiv.txt

Black holes in General Relativity are very simple objects. This property, that goes under the name of ``no-hair,'' has been refined in the last few decades and admits several versions. The simplicity of black holes makes them ideal testbeds of fundamental physics and of General Relativity itself. Here we discuss the no-hair property of black holes, how it can be measured in the electromagnetic or gravitational window, and what it can possibly tell us about our universe.

the no-hair hypothesis (and why should we care)?} \subsection{The simplicity of black holes in General Relativity\label{sec:Sch}} We celebrate this year the first direct detection of gravitational waves (GWs) and the first detection of a black hole (BH) binary, in its last stages of coalescence~\cite{Abbott:2016blz}. In this context, it is appropriate to also honor the centenary of the Schwarzschild solution, which describes any regular asymptotically flat, static and spherically symmetric vacuum spacetime in General Relativity (GR). In standard Schwarzschild coordinates, the solution reads \be ds^2=-c^2\left(1-\frac{2GM}{c^2r}\right)dt^2+\left(1-\frac{2GM}{c^2r}\right)^{-1}dr^2+r^2d\Omega^2\,,\label{metric_Schwarzschild} \ee where $G$ is Newton's constant and $c$ the speed of light. The most distinctive feature of the solution above is a coordinate singularity at $r=2GM/c^2$, describing a null surface (the event horizon) which causally separates the inside and outside regions. For physical setups where (static, spherically symmetric) matter is present, the solution should be truncated at the radius of the object and merged smoothly with some interior solution. For pure vacuum spacetimes, it describes a BH. In other words, any static BH in the Universe which is non-spinning and that lives in approximately empty surroundings is described by the geometry~(\ref{metric_Schwarzschild}). The Schwarzschild solution is fully characterized by a single parameter, the total gravitational mass $M$. In this respect, it is not dissimilar from its Newtonian counterpart: a spherically symmetric, vacuum, static solution of Newton's gravity is also described by only a mass parameter. It turns out to be difficult to construct BH solutions described by more parameters. For example, let's try to ``anchor'' a weak, {\it static} massless spin-$s$ field onto the Schwarzschild solution. For the sake of illustration we focus on minimally coupled scalars, vector fields described by Maxwell's theory and gravitational fluctuations within vacuum GR~\cite{Berti:2009kk}. At a linearized level (i.e., keeping the background geometry fixed), these fields can be expanded in scalar, vector or tensor harmonic functions, parametrized by an integer number $l=0,1,2,...$, which carry information on the angular dependence of the field. These fields are all described by the equation~\cite{Berti:2009kk} \be \left(\left(1-\frac{2GM}{c^2r}\right) \Psi'\right)'-\left(\frac{l(l+1)}{r^2}+\frac{2GM(1-s^2)}{c^2r^3}\right)\Psi=0\,, \ee where $\Psi$ denotes the field amplitude, $s=0,1,2$ for scalars, vectors and tensors, respectively, and primes stand for radial derivatives. One can now multiply the above equation by the complex conjugate $\Psi^*$, integrate from the horizon to infinity, and look for regular solutions of the above equation. We get, upon performing an integration by parts and dropping a boundary term (regular solutions evaluate it to zero), \be -\int_{\frac{2GM}{c^2}}^{+\infty}dr\left(1-\frac{2GM}{c^2r}\right)|\Psi'|^2+\frac{l(l+1)}{r^3}\left(r+\frac{2GM(1-s^2)}{c^2\,l(l+1)}\right)|\Psi|^2=0\,. \ee For any $l\geq s$ the integrand is negative-definite outside the horizon, and the only solution to the above equation is the trivial one, $\Psi=0$. For vector fields, there is a nontrivial solution $\Psi={\rm const}$ for $l=0$, describing a weakly charged BH. For tensors, the only non-trivial solutions (for which the integrand is not positive-definite) have $l=0, 1$ and correspond to a slight change of mass and addition of small amount of

It is by now well-known that the Kerr family is {\it not} the most general solution of Einstein's field equations in the presence of reasonable forms of matter. However, there are good reasons to believe that BHs which do not belong to the Kerr family are either dynamically unstable or do not form out of realistic collapse scenarios. The no-hair or Kerr hypothesis is therefore a cherished belief, which upcoming experiments can test, either in the GW or electromagnetic band. Some of these measurements will be more sensitive to the null geodesic around compact objects, than to the presence of the event horizon itself~\cite{Cardoso:2016rao}... but given that even light rings are a unique feature of relativistic theories of gravity, these are all exciting years ahead. \vspace{1cm} \noindent {\bf Acknowledgments.} We thank the anonymous referee for useful suggestions, which contributed to improve the quality of the manuscript. We are indebted to Jo\~ao Costa, Carlos Herdeiro, Chris Moore, Thomas Sotiriou, Norbert Wex, and Nicolas Yunes for useful comments on the manuscript. V. C thanks the Physics Department of the University of Rome ``La Sapienza'' for hospitality while this work was being completed. V.C. acknowledges financial support provided under the European Union's H2020 ERC Consolidator Grant ``Matter and strong-field gravity: New frontiers in Einstein's theory'' grant agreement no. MaGRaTh--646597, and FCT for Sabbatical Fellowship nr. SFRH/BSAB/105955/2014. Research at Perimeter Institute is supported by the Government of Canada through Industry Canada and by the Province of Ontario through the Ministry of Economic Development $\&$ Innovation. This work was supported by the H2020-MSCA-RISE-2015 Grant No. StronGrHEP-690904. \vskip 1cm \addcontentsline{toc}{section}{Appendix: Multipole moments in Newtonian gravity and in general relativity}

1607.03133

1607

1607.00306_arXiv.txt

Using high-resolution transition region (TR) observations taken by the Interface Region Imaging Spectrograph (IRIS) mission, \citet{Tian+etal.BD.2014ApJ...790L..29T} revealed numerous short-lived subarcsecond bright dots (BDs) above sunspots (mostly located in the penumbrae), which indicate yet unexplained small-scale energy releases. Moreover, whether these subarcsecond TR brightenings have any signature in the lower atmosphere and how they are formed are still not fully resolved. This paper presents a multi-wavelength study of the TR penumbral BDs using a coordinated observation of a near disk-center sunspot with IRIS and the 1.6~m New Solar Telescope (NST) at the Big Bear Solar Observatory. NST provides high-resolution chromospheric and photospheric observations with narrow-band \ha\ imaging spectroscopy and broad-band TiO images, respectively, complementary to IRIS TR observations. A total of 2692 TR penumbral BDs are identified from a 37-minute time series of IRIS 1400~\AA\ slitjaw images. Their locations tend to be associated more with downflowing and darker fibrils in the chromosphere, and weakly associated with bright penumbral features in the photosphere. However, temporal evolution analyses of the BDs show that there is no consistent and convincing brightening response in the chromosphere. These results are compatible with a formation mechanism of the TR penumbral BDs by falling plasma from coronal heights along more vertical and dense magnetic loops. The BDs may also be produced by small-scale impulsive magnetic reconnection taking place sufficiently high in the atmosphere that has no energy release in the chromosphere.

While sunspots have been observed for a long time, their fine structure and vertical extension are only revealed in recent decades by high-resolution and multi-wavelength (beyond visible) observations. Sunspots are concentrations of strong magnetic fields emerged from the convection zone. The magnetic fields are more vertical in the umbra and more inclined in the penumbra. In the photosphere, penumbral magnetic fields are observed to consist of two components, i.e., a more horizontal component carrying most of the outward Evershed flows is embedded in a more vertical background component \citep[the so-called ``uncombed'' or ``interlocking-comb'' structure, e.g., ][]{Solanki+Montavon1993A&A...275..283S, Lites+etal1993ApJ...418..928L, BellotRubio+etal2004A&A...427..319B, Ichimoto+etal2007PASJ...59S.593I, Deng+etal2010ApJ...719..385D, Borrero+Ichimoto2011LRSP....8....4B}. Recent high-resolution observations \citep[e.g.,][]{Ichimoto2010mcia.conf..186I} and realistic three-dimensional radiative magnetohydrodynamic simulations \citep{Rempel+etal2009Sci...325..171R, Rempel+etal2009ApJ...691..640R, Kitiashvili+etal2009ApJ...700L.178K} show that the thermal magnetoconvection is responsible for the observed Evershed effect and the filamentary intensity structures in the penumbrae. The bright penumbral grains/fibrils are locations of up-flowing hot gas associated with more vertical magnetic fields \citep[e.g.,][]{Langhans+etal2005A&A...436.1087L, rimmele+marino2006, Wang+Deng+Liu2012ApJ...748...76W}. The dark penumbral filaments are locations corresponding to cooled horizontal Evershed flows guided by more horizontal magnetic fields \citep[e.g.,][]{stanchfield+thomas+lites1997, Rempel2011ApJ...729....5R}. In the chromosphere, the filamentary penumbral structures are still present and extend out to a larger area called the superpenumbra. Inverse (i.e., inward) Evershed flow is observed in the chromospheric penumbra and superpenumbra \citep[][and references therein]{Solanki2003A&ARv..11..153S}. It is found that the penumbral filamentary structures seen at sufficiently different heights (for example, the formation heights of a strong photospheric line core and the continuum, which are a few hundreds of kilometers apart) are poorly correlated \citep{Wiehr+Stellmacher1989A&A...225..528W}, implying limited vertical extension for those features. Besides the aforementioned long lasting and relatively stable features (usually lasting more than 20 minutes), small-scale and short lived dynamic transients are also observed in the penumbral atmosphere. Using high-resolution chromospheric \CaIIH\ filtergrams obtained by the Solar Optical Telescope \citep[SOT;][]{Tsuneta+etal2008SoPh..249..167T} onboard Hinode, \citet{Katsukawa+etal2007Sci...318.1594K} reported ``penumbral microjets'' that are small-scale jet-like brightening features (width $\sim$400 km, length 1000--4000 km, duration $<$ 1 min) ubiquitously and constantly present above sunspot penumbrae. They are found to emanate from locations near the bright penumbral grains in between dark penumbral filaments. The authors argued that the penumbral microjets might be generated by magnetic reconnection due to the drifting motion of penumbral grains in the uncombed magnetic configuration. \citet{Reardon+etal2013ApJ...779..143R} studied the spectral properties of penumbral transient brightenings in the upper photosphere and chromosphere using \CaII\ 854.21~nm spectroscopy obtained by the Interferometric Bidimensional Spectrometer \citep[IBIS;][]{Cavallini2006SoPh..236..415C}. They found that penumbral microjets show double-humped spectral line profiles (i.e., emission in both wings) that are similar to those of Ellerman bombs \citep{ellerman17}. Moreover, they identified other types of transient brightenings that show significantly different spectral characteristics, such as acoustic shocks, umbral flashes, and ``chromospheric modification". They concluded that the apparent highly dynamic brightenings above sunspots can be produced by different physical mechanisms. The gas in the transition region (TR) above sunspots is even more complex and dynamic \citep{Solanki2003A&ARv..11..153S}. The recently launched Interface Region Imaging Spectrograph \citep[IRIS;][]{DePontieu.IRIS.2014SoPh..289.2733D} mission provides unprecedented observations of the solar TR and chromosphere with high cadence (several seconds), high spatial (0\arcsec.167~pixel$^{-1}$) and spectral resolution. Using IRIS 1400~\AA\ and 1330~\AA\ slit-jaw images (SJIs), \citet{Tian+etal.BD.2014ApJ...790L..29T} revealed numerous short-lived subarcsecond bright dots (BDs) in the TR above every sunspot they investigated (located mainly in the penumbrae). These TR penumbral BDs often have intensities a few times stronger than their surrounding background in the 1400~\AA\ SJIs. They generally appear slightly elongated along the sunspot radial direction with typical sizes of a few hundreds of kilometers. Despite some long-lasting (several minutes or even longer) strong dots, most of the BDs last no more than a minute. When a BD occurs, the TR line profiles (e.g., \SiIV\ 1402.77 \AA\ and \CII\ 1334.53 \AA) are strongly enhanced and broadened, suggesting strong emission in the TR. Some strong (i.e., very bright, relatively large and long-lasting) dots can also be seen in the 304 \AA, 171 \AA, 193 \AA, 211 \AA, and 131 \AA\ passbands of Atmospheric Imaging Assembly \citep[AIA;][]{Lemen+etal2012SoPh..275...17L} onboard Solar Dynamics Observatory \citep[SDO;][]{Pesnell.SDO.2012SoPh..275....3P}. They appear at the TR footpoints of coronal loops and are likely related to small-scale energy release events there. The thermal energy of the BDs is estimated to be in the order of 10$^{22}$--10$^{23}$ erg, which is in the scope of nanoflares \citep{ParkerNanoFlare1988ApJ...330..474P}. Using High-resolution Coronal imager (Hi-C) 193 \AA\ images, \citet{Alpert+etal2016ApJ...822...35A} also observed BDs in a sunspot penumbra and performed a similar analysis to characterize their physical properties. Hi-C penumbral BDs are on average slower, dimmer, larger in size and longer lived than IRIS penumbral BDs. Analysis of light curves of different AIA passbands suggests that the temperatures of most Hi-C penumbral BDs are likely in the TR range. It is, however, not clear yet whether these TR penumbral BDs have any signatures in the lower atmosphere and how are they formed. \citet{Tian+etal.BD.2014ApJ...790L..29T} proposed two possible generation mechanisms for them: 1. small-scale magnetic reconnection in the TR and chromosphere involving the uncombed penumbral magnetic fields; 2. associated with falling plasma. To fully understand the origin of the TR penumbral subarcsecond BDs and their relationship with the aforementioned lower-atmospheric penumbral features, co-temporal observations in the chromosphere and photosphere with high spatial and temporal resolution are necessary. This paper presents a joint study of these BDs using a coordinated observation by IRIS and the 1.6~m New Solar Telescope \citep[NST;][]{GoodeNST2010AN....331..620G} at the Big Bear Solar Observatory (BBSO). The \ha\ imaging spectroscopy and the TiO images from BBSO/NST provide important information in the chromosphere and photosphere, respectively, which enables us to investigate whether and to what extent a correspondence of the TR penumbral BDs exists in the lower atmosphere.

\label{sec:summary} We have presented a statistical study of TR penumbral subarcsecond BDs using multi-wavelength observations from IRIS and NST. A total of 2692 BDs are automatically identified from a 37-minute time series of IRIS 1400~\AA\ SJIs. They have a mean size of $\sim$400~km and a mean lifetime of $\sim$43~s. These statistical properties are very similar to those presented in \citet{Tian+etal.BD.2014ApJ...790L..29T}. The locations of TR penumbral BDs tend to be associated more with redshifted (i.e., downflowing) and darker features in the chromosphere, and weakly associated with brighter features in the photosphere. Analyses of multi-wavelength temporal light curves of the BDs show that the brightening response dramatically decreases with depth. Only weak brightening response is found in the upper chromosphere. There is no noticeable and consistent response below the middle chromosphere. Our results exclude the possibility that the TR penumbral BDs are directly generated by magnetic reconnection taking place in the chromosphere, because no convincing energy release signature was found in the chromosphere. It is still possible that the small-scale magnetic reconnection occurs sufficiently high in the atmosphere, i.e., in the middle TR or above, which causes localized impulsive energy release. The falling plasma scenario is compatible with our observational results, therefore it is also very likely. Subarcsecond heating/brightening events are recently observed in the TR or corona in different places by IRIS and Hi-C. For example, using IRIS observations, \citet{Kleint+etal2014ApJ...789L..42K} found small-scale heating events associated with supersonic downflows in the TR above sunspot umbrae and explained them with falling plasma mechanism. \citet{Skogsrud+etal2016ApJ...817..124S} studied bright roundish small patches, so-called bright ``grains", which ubiquitously appear in IRIS 1400 channel in active region plages. The bright grains have sizes of 0\arcsec.5 -- 1\arcsec.7 and lifetimes of a few minutes. Many of them show good correspondence with the behavior of shock-driven ``dynamic fibrils" observed in \ha. The authors thus suggest that the bright grains are the result of chromospheric shocks impacting the TR. Using Hi-C observation, \citet{Regnier+etal2014ApJ...784..134R} reported sparkling extreme-ultraviolet (EUV) BDs at the edge of an active region. These EUV BDs have similar sizes and lifetimes as the TR penumbral BDs, although they occur at a different place. They are interpreted as small-scale impulsive heating events at the base of coronal loops. Combining observations from Swedish 1-m Solar Telescope (SST) in \CaII\ lines and IRIS for a near limb sunspot, \citet{Vissers+etal2015ApJ...811L..33V} found evidence for a TR brightening response to chromospheric penumbral microjets along the jet extent. In contrast to our analysis, they used an opposite approach. They first identify chromospheric penumbral bright jets using SST \CaII\ data, then seek TR response signatures using IRIS data. Considering the similar sizes and lifetimes between the penumbral microjets and the TR penumbral BDs, the authors speculate that many moving TR BDs are the tops of these jets. We cannot exclude that some TR BDs might be related to the chromospheric penumbral bright jets, considering the weak emission response in 2796~\AA\ and a small fraction showing positive correlation with \ha\ center light curves. However, they could not be the majority. \citet{Tian+etal.BD.2014ApJ...790L..29T} reported that about 32\% of BDs move inward and only 13\% move outward. The jet top scenario maybe more suitable for those outward moving BDs which only account for a small fraction. The inward moving BDs seem favor more falling motion, rather than jetting motion. Using Hinode/SOT, Hi-C, and SDO/AIA observations of a sunspot, \citet{Tiwari+etal2016ApJ...816...92T} also examined whether chromospheric penumbral microjets have signatures in the TR and corona. They found that most penumbral microjets hardly have any discernible signature in most AIA passbands except for some exceptionally stronger/larger jets occurred in the tails of some penumbral filaments where opposite-polarity magnetic fields are seen. These tail penumbral jets do have brightening signatures in the TR. Future development of high-resolution magnetic field measurements in the chromosphere \citep[e.g.,][]{delaCruz+Socas-Navarro2011A&A...527L...8D, Rouppe+delaCruz2013ApJ...776...56R, Schad+etal2013ApJ...768..111S} will allow us to directly investigate the relation of BDs to the magnetic field topology of the lower solar atmosphere.

1607.00306

1607

1607.07710_arXiv.txt

We present $\simeq0\farcs4$-resolution extinction-independent distributions of star formation and dust in 11 star-forming galaxies (SFGs) at $z = 1.3-3.0$. These galaxies are selected from sensitive, blank-field surveys of the $2' \times 2'$ Hubble Ultra-Deep Field at $\lambda = 5$ cm and 1.3 mm using the Karl G. Jansky Very Large Array (VLA) and Atacama Large Millimeter/submillimeter Array (ALMA). They have star-formation rates (SFRs), stellar masses, and dust properties representative of massive main-sequence SFGs at $z \sim 2$. Morphological classification performed on spatially-resolved stellar mass maps indicates a mixture of disk and morphologically disturbed systems; half of the sample harbor X-ray active galactic nuclei (AGN), thereby representing a diversity of $z \sim 2$ SFGs undergoing vigorous mass assembly. We find that their intense star formation most frequently occurs at the location of stellar-mass concentration and extends over an area comparable to their stellar-mass distribution, with a median diameter of $4.2 \pm 1.8$ kpc. This provides direct evidence for galaxy-wide star formation in distant, blank-field-selected main-sequence SFGs. The typical galactic-average SFR surface density is 2.5 M$_{\sun}$yr$^{-1}$kpc$^{-2}$, sufficiently high to drive outflows. In X-ray-selected AGN where radio emission is enhanced over the level associated with star formation, the radio excess pinpoints the AGN, which are found to be co-spatial with star formation. The median extinction-independent size of main-sequence SFGs is two times larger than those of bright submillimeter galaxies whose SFRs are $3-8$ times larger, providing a constraint on the characteristic SFR ($\sim300$ M$_{\sun}$yr$^{-1}$) above which a significant population of more compact star-forming galaxies appears to emerge.

\label{sec:intro} Numerical simulations and observational inferences suggest that typical star-forming galaxies (SFGs) at the peak of galaxy assembly activity, $z \simeq 1 - 3$, assembled most of their stellar mass via accretion of cold gas, which led to gas-rich, unstable disks and disk-wide star formation \citep[e.g.,][]{Keres05, Dekel09, Bournaud09}. The isolated, in-situ assembly of typical SFGs is inferred from the presence of a relationship between star formation and stellar mass, the ``main sequence'' \citep[e.g.,][]{Noeske07, Whitaker12, Speagle14}, and from the rarity of compact starbursts at $z \sim 2$, as indicated by their specific star-formation rate (SFR) distribution and infrared color \citep{Rodighiero11, Elbaz11}. At higher SFRs ($\gtrsim300$ \Msunyr\ at $z \sim 2$), there is no theoretical consensus on whether mergers or continuous accretion are the dominant triggering mechanism of intense star formation \citep{Hopkins10, Dave10, Hayward13, Narayanan15}; the relative contribution of in-situ versus merger modes at this SFR regime have not been constrained observationally. Contrary to the inferred in-situ assembly of SFGs, {\it Hubble Space Telescope} ({\it HST}) observations of main-sequence SFGs at $z \sim 2$ commonly show galactic substructures and asymmetry that are signposts of mergers that drive intense star formation in the local Universe \citep{Lotz04, Kartaltepe12}. However, some substructures, such as optically-bright star-forming clumps with SFR $\simeq1 - 30$ of \Msunyr\ \citep[e.g.,][]{FS09, Guo15} can also be a natural consequence of gas-rich, turbulent disks evolving in isolation. Yet the role of star-forming clumps in assembling the bulk of stellar mass is debated \citep{Genel12, Bournaud14} and optically-selected clumps altogether contain $< 10-20\%$ of the total star formation of their host \citep{Wuyts12, Guo12, Guo15}. Directly observing the distribution of the bulk of star formation in galaxies at $z \sim 2$ is therefore key to establishing the relative contribution between modes of star and bulge formation. Significant progress in directly imaging the distribution of star formation at $z > 1$ has been made with spatially-resolved H$\alpha$ spectroscopy, e.g., SINS \citep{FS09}, KMOS$^{\rm 3D}$ \citep{Wisnioski15}, and KROSS \citep{Stott16}. Average H$\alpha$ maps from 3D-HST, the largest sample thus far of resolved star formation at $z = 1.5 - 2.5$, show that the star-formation surface density, \SFRSD, on average peaks near the centers of massive galaxies \citep{Nelson15}. However, the dust extinction also peaks at the center, such that a factor of 6 correction to the inferred H$\alpha$ SFR is required in the central kpc \citep[and more than a factor of 10 close to the center;][]{Nelson16}. Above SFRs as small as 20 \Msunyr, which is $0.2\times$ the typical rate for main-sequence SFGs at $z > 1$, galaxies become so dust enshrouded that almost no light emerges in rest-frame ultraviolet observations \citep{Reddy10}. The question of where exactly new stars form within typical $z \sim 2$ SFGs is hence a deceptively simple one that is challenging to address. Breakthroughs in this area require sub-arcsecond resolution, extinction-independent star-formation tracers for main-sequence SFGs at $z > 1$, which are now available with the Karl G. Jansky Very Large Array (VLA) and the Atacama Large Millimeter/submillimeter Array (ALMA). To this end, we conduct two sensitive, blank-field imaging surveys of the Hubble Ultra-Deep Field (HUDF, $\alpha = 03^{\rm h}$32$^{\rm m}$, $\delta =-27^{\circ}47'$) using the VLA and ALMA at $\lambda = 5$ cm and 1.3 mm, respectively, to make $0\farcs4$ resolution images of SFGs at $z \sim 2$. The 5 cm continuum traces star formation through the synchrotron emission from supernova remnants, but can be affected by AGN emission; whereas the 1.3 mm continuum traces cold dust associated with star formation, but requires uncertain assumptions about the shapes of spectral energy distributions (SEDs) to estimate the SFR. The combination of the two surveys therefore provides complementary strengths, especially in the HUDF where the wealth of ancillary data can help, e.g., identify AGN. By establishing that the VLA and ALMA trace the common extent of star formation, ALMA can serve as a morphological tracer of ``pure'' star formation in AGN hosts because the 1.3-mm dust continuum is neither contaminated by AGN torus emission \citep{Elvis94, Mullaney11, MorNetzer12} nor synchrotron emission from the jets\footnote{Assuming a flat spectral index for AGN, the radio continuum from the brightest radio AGN in the HUDF is less than 1\% of their 1.3 mm emission from star formation.}. Furthermore, because ALMA significantly gains in sensitivity to star formation with the help of negative K-corrections at $z > 2.5$, whereas the VLA gradually loses star-formation sensitivity beyond this redshift, the combination of VLA and ALMA yields a sensitive probe of the morphology of star formation over the entire range of $z = 1 - 3$. In this paper, we present first results from combining the VLA and ALMA HUDF surveys, focusing on the extinction-independent distributions of the star formation in SFGs at $z = 1-3$. We discuss the VLA and ALMA surveys and ancillary data in \S\ref{sec:observations} and present the size and location of star-formation in SFGs selected from VLA and ALMA, along with their implications in \S\ref{sec:results}. We adopt a $\Lambda$CDM cosmology with $\Omega_M = 0.3$, $\Omega_\Lambda = 0.7$, $H_{0} = 70~{\rm km\,s}^{-1}{\rm Mpc}^{-1}$, and the \citet{Chabrier03} IMF. \begin{figure*}[t] \figurenum{1} \centerline{\includegraphics[width=0.85\textwidth]{Fig1_1of2.pdf}} \caption{The 11 SFGs detected in both the VLA and ALMA HUDF surveys. From left to right: {\it HST} $i_{814}-J_{125}-H_{160}$ color composite; unobscured star formation ({\it HST}/F606W); rest-frame optical morphology ({\it HST}/F160W); star-formation rate surface density, \SFRSD\ (VLA, details in \S\ref{sec:observations}); dust mass surface density, \DustSD\ (ALMA, details in \S\ref{sec:gas}); and stellar-mass surface density, \SigmaMstar\ (from spatially-resolved SED fitting using {\it HST} images). Each image is $4'' \times 4''$; North is up, East is on the left. VLA and ALMA synthesized beams are shown in the corresponding columns; the contours are $[-3, 3^{1}, 3^{1.5}, 3^{2}, ...] \times \sigma$ for VLA and $[-2.5, 2.5^{1}, 2.5^{1.5}, 2.5^{2}, ...] \times \sigma$ for ALMA; negative contours are shown as dotted lines. Intense star formation most frequently occurs within the stellar-mass concentration and extends over a large area of the stellar-mass buildup, i.e., galactic-wide (continued on next page). \label{fig:sixpanA}} \end{figure*} \begin{figure*}[t] \figurenum{1} \centerline{\includegraphics[width=0.85\textwidth]{Fig1_2of2.pdf}} \caption{(continued) We note that the radio \SFRSD\ maps for UDF7, 11, and 13 that harbor X-ray-selected AGN with radio emission enhanced more than twice the level of star-formation emission (Figure \ref{fig:FRC}, tabulated in Table \ref{tab:sourcetable}) may contain significant contributions from the AGN and hence marked as ``radio AGN''.} \end{figure*}

\label{conclude} We have made the first comparison between ultra-deep VLA and ALMA imaging of $z \sim 2$ main-sequence SFGs. The far-infrared/radio correlation appears to hold for individual main-sequence SFGs with SFR $\sim$ 100 \Msunyr\ at $z \sim 2$ and the extinction-independent distributions of star formation are consistent between the data sets. The intense star formation in our blank-field-selected sample extends over a large galactic area regardless of their stellar-mass morphology (isolated or morphologically disturbed) or the presence of AGN, with the SFR and dust-mass surface densities both peaking near the existing stellar-mass concentration. These findings provide direct-imaging evidence of a gas-rich galactic environment with widespread occurrence of intense star formation. The spatially-resolved SFR surface densities are sufficiently large across the areas of dominant stellar-mass buildups that they may drive galaxy-wide outflows. Where radio excess permits pinpointing of the AGN, it is found to be co-spatial with the dust mass concentrations. The median star-formation diameter in our main-sequence SFGs sample is $4.2 \pm 1.8$ kpc, two times larger than those of sub-mm galaxies forming stars at $3-8$ times higher than the main-sequence SFGs, indicating that, at $z \sim 2$, the SFR threshold above which a significant population of more compact SFGs appears to emerge is $\sim300$ \Msunyr.

1607.07710

1607

1607.05715_arXiv.txt

We discuss dynamical systems approaches and methods applied to flat Robertson-Walker models in $f(R)$-gravity. We argue that a complete description of the solution space of a model requires a global state space analysis that motivates globally covering state space adapted variables. This is shown explicitly by an illustrative example, $f(R) = R + \alpha R^2$, $\alpha > 0$, for which we introduce new regular dynamical systems on global compactly extended state spaces for the Jordan and Einstein frames. This example also allows us to illustrate several local and global dynamical systems techniques involving, e.g., blow ups of nilpotent fixed points, center manifold analysis, averaging, and use of monotone functions. As a result of applying dynamical systems methods to globally state space adapted dynamical systems formulations, we obtain pictures of the entire solution spaces in both the Jordan and the Einstein frames. This shows, e.g., that due to the domain of the conformal transformation between the Jordan and Einstein frames, not all the solutions in the Jordan frame are completely contained in the Einstein frame. We also make comparisons with previous dynamical systems approaches to $f(R)$ cosmology and discuss their advantages and disadvantages.

The simplest class of fourth order metric gravity theories is based on an action \begin{equation}\label{actionJ} \mathcal{S} = \int\left\{\frac{f(R)}{2\kappa^2} + \mathcal{L}_{m} \right\} \sqrt{-\det{g}}\, d^{4}x \end{equation} where $\kappa^2 = 8\pi G$; the speed of light, $c$, is set to one; $\det{g}$ is the determinant of a Lorentzian 4-dimensional metric $g$, and $R$ the associated curvature scalar, while $\mathcal{L}_{m}$ is the matter Lagrangian density. General relativity with a cosmological constant $\Lambda$ is obtained by setting $f(R)=R-2\Lambda$. The vacuum part of these models, i.e., $\mathcal{L}_{m}=0$, has recently achieved some popularity where certain forms of the function $f(R)$ have resulted in geometric models of inflation or, more recently, dark energy, see e.g.~\cite{capfra08}--\cite{aveetal16} and also~\cite{sch07,barott83} for a historical background. Although an assessment of cosmological viability requires a study of spatially homogeneous and isotropic Robertson-Walker (RW) models and perturbations thereof, we will restrict the analysis in this paper to flat RW cosmology. The vacuum equations of these models can be written as (see e.g.~\cite{LivRev10}): \begin{subequations}\label{Jordanfulleq} \begin{align} \dot{a} &= Ha, \label{aeqf}\\ \dot{H} &= -2H^2 + \frac{R}{6}, \label{dotHf}\\ \ddot{R} &= -3H\dot{R} - \frac{1}{F_{,R}}\left[F_{,RR}\dot{R}^2 + \frac13\left(2f - FR\right) \right], \label{ddotRf} \end{align} \begin{equation} -6H\left(F_{,R}\dot{R} + FH\right) + FR - f = 0 .\label{constJorf} \end{equation} \end{subequations} In the above equations an overdot represents the Jordan proper time $t$ derivative, $a$ is the scale factor of the flat RW metric in the Jordan frame, $H$ is the Jordan Hubble variable, $R$ is the Jordan curvature scalar, and \begin{equation} F = \frac{df}{dR}, \qquad F_{,R} = \frac{dF}{dR} = \frac{d^2f}{dR^2}, \qquad F_{,RR} = \frac{d^2F}{dR^2} = \frac{d^3f}{dR^3}. \end{equation} By regarding $\dot{R}$ as an independent variable, equation~\eqref{constJorf} forms a constraint that the evolution equations must satisfy. Furthermore, equation~\eqref{aeqf} for $a$ decouples, which leads to a reduced closed system of first order equations for $(H,\dot{R},R)$, which, due to the constraint~\eqref{constJorf}, yield a dynamical system describing a flow on a 2-dimensional state space. Once the reduced system of first order equations has been solved, the decoupled equation~\eqref{aeqf} yields $a \propto \exp(\int dt H $). The above system of equations has some general properties which are worth mentioning. Firstly, the system is invariant under the transformation $(t,H) \rightarrow -(t,H)$, a property we will use below. Secondly, the system is ill-defined whenever $F_{,R}=0$ for some value(s) of $R$.\footnote{Of course this is not the case for general relativity for which $F_{,R}$ is identically zero, a case we will not consider here.} This is related to pathological properties as regards the characteristics of $f(R)$ gravity, where, e.g., the properties of gravitational waves will severely constrain the physical viability of such models. It is therefore natural to divide $f(R)$ gravity into two main classes of models: those for which $F_{,R}>0$ and those for which this is not the case. Thirdly, as it is well known $F>0$ is associated with that one can introduce an Einstein frame. However, $F=0$ is not, in general, an invariant subset in the Jordan frame, since \begin{equation}\label{dotFgen} \dot{F}|_{F=0} = F_{,R}\dot{R}|_{F=0} = - \left. \frac{f}{6H}\right|_{F=0}, \end{equation} where we have used~\eqref{constJorf} (i.e., $F=0$ is only an invariant subset if $f$ and $F$ are simultaneously zero for some value of $R$). This suggest that it is also natural to divide $f(R)$ models into two additional classes, those with $F>0$ everywhere, and those for which $F$ can change sign. The latter case yields solutions in the Einstein frame that can be conformally extended in the Jordan frame, but which ones depend on the explicit form of $f(R)$. % We will later explicitly illustrate this result in the context of a specific model, which also exemplifies some other general features of $f(R)$ cosmology. There are a number of dynamical systems formulations in the literature that are based on transformations from $(H,\dot{R},R)$ to some other variables (see Appendix~\ref{app:systems} for a discussion on several of these formulations). In this context there has been considerable activity concerning fixed points (also called singular points, equilibrium points, critical points) and their linear stability properties. It is therefore of interest to consider the fixed points of the basic reduced state space variables $(H,\dot{R},R)$, which must satisfy \begin{equation}\label{fixedgen} -2H^2 + \frac{R}{6} = 0 ,\qquad \dot{R} = 0, \qquad 2f - FR =0, \end{equation} as follows from~\eqref{dotHf} and~\eqref{ddotRf}. Then,~\eqref{constJorf} takes the form (assuming that $F_{,R}$ is non-zero) \begin{equation} -6H^2F + FR - f = 3F(-2H^2 + R/6)=0, \end{equation} which is thereby automatically satisfied. Moreover, $2f - FR$ is identically zero, if and only if $f \propto R^2$, and only in this case there is a line of fixed points for which $R = 12H^2$, while all other models have isolated fixed points. When transforming to other variables care has to be taken when it comes to the physical interpretation of fixed point results. As we will see, some fixed points in other formulations simply reflect a break down of those variables, i.e., they correspond to a state space coordinate singularity. It is also essential to note that fixed points do not always give a complete asymptotic description. As will be emphasized in this paper, it is necessary to consider the dynamics on the entire state space of a given model to make an assessment of its physical content. What is then required in order to obtain a complete description of the solution space and the properties of those solutions for a given $f(R)$ RW model? We will illustrate some of the ingredients that are required to answer this question with a specific example, but for all $f(R)$ RW cosmologies one needs to do the following: \begin{itemize} \item[(i)] State space analysis. \item[(ii)] A complete state space adapted coordinate cover, including those state space boundaries for which the equations can be extended (this e.g. excludes boundaries for which $F_{,R}$ becomes zero). \item[(iii)] Local and global dynamical systems analysis. \item[(iv)] Physical solution space interpretation. \end{itemize} Let us now comment on the above in a little more detail. (i) A state space analysis entails dimensional and scale considerations, and a study of the algebraic structure of the constraint equation~\eqref{constJorf}, which includes global aspects such as state space topology. (ii) This means that one needs to find state space coordinates that \emph{globally} cover the state space of a given model, including the boundaries for which the equations can be differentiable extended. This may include limits where $H$, $\dot{R}$ and $R$ become unbounded, which motivates the introduction of new \emph{bounded} variables. Note that some models will even in principle require several coordinate patches, but there are classes of models for which one can find common \emph{local} useful variables. Even in cases where it is possible to find a bounded global state space coordinate system, it might still be useful to consider other variables since it is unlikely that a global system, except under very special circumstances, is the optimal one for all local structures, i.e., there might exist complementary sets of variables. Furthermore, different models have different state space structures, and in general this requires different choices of variables --- the common element is instead a state space analysis and an adaption to the structures that the analysis reveals. (iii) To understand the solution space structure of a given model and the asymptotical behaviour of the solutions, which is essential for assessing its physical viability (it is not enough to consider special solutions, e.g., fixed points), one must, in general, apply linear \emph{and} non-linear fixed point techniques, as well as \emph{global} dynamical systems analysis. Furthermore, note that fixed points will not in general give a complete asymptotic description, e.g., a problem might naturally give rise to limit cycles. That a global understanding of the solution space is required is illustrated by the fact that if one has found a solution with a desirable evolution, then the models will still only be of interest if this solution is in some sense an `attractor solution.' Even so, this does not exclude that there exists an additional set of solutions that have a different evolution, which leads to issues concerning measures describing how `typical' a solution is. (iv) Solutions, e.g. fixed points, have to be physically interpreted since a solution might be an artifact of the variables one has used. For example, variables that do not cover the entire Jordan state space result in coordinate singularities, which results in fixed points. Thus fixed points may not correspond to physical phenomena, but may instead show that a formulation breaks down. To illustrate the above issues (excluding the situation where $F_{,R}$ passes through zero, which we will comment on in the final discussion), as well as allowing us to introduce some dynamical systems methods of quite wide applicability, we will consider a specific example, the vacuum equations for the flat RW metric with \begin{equation}\label{fR2} f(R) = R + \alpha R^2, \qquad \alpha >0 . \end{equation} This model has attracted considerable attention in the past, see e.g.,~\cite{Sta80}--\cite{gen15}, and it still remains as one of the more successful models of inflation~\cite{plaXX15}. Although some interesting results have been obtained, previous analyses have been severely hampered by formulations that do not give a complete, or sometimes correct, description of the global solution space and its properties. In contrast, we will here give a complete description of the entire solution space of these models, and we will also describe the solutions' asymptotic behaviour. More importantly though is that this model allows us to explicitly address some aspects about how to obtain useful dynamical systems treatments of RW $f(R)$ models, and to illustrate various dynamical systems methods. For example, we will situate the entire solution space of the Einstein frame in the state space of the Jordan frame, which allows us to explicitly show how some solutions in the Einstein frame for these models correspond to entire solutions in the Jordan frame, while other solutions can be conformally extended in the Jordan frame. In other words, a local dictionary between the two frames does not always entail global equivalence (the curious reader can skip ahead and take a look at Figures~\ref{fig:VacuumJF}, \ref{fig:Einsteinstatespace} and~\ref{fig:Jordan_EinsteinBound} below). The outline of the paper is as follows. In the next section we make a state space analysis for the $f(R) = R + \alpha R^2$ models, which is used to produce a new regular unconstrained dynamical system formulation on a compact state space for the Jordan frame. We then use this system to perform a local analysis of the fixed points, focusing on non-linear aspects such as blow ups of nilpotent fixed points. This is followed by a global analysis that gives a complete description of the entire solution space of the models, which is depicted and summarized in Figure~\ref{fig:VacuumJF}. We emphasize the importance of the global topological structure of the state space for a full understanding of the solution space. In Section~\ref{sec:ein} we present a new regular unconstrained dynamical system formulation on a compact state space for the Einstein frame. We then perform a local analysis of fixed points, again focusing on non-linear aspects such as center manifold analysis. It is also shown that the breakdown of the Einstein frame variables at $F=0$ leads to fixed points in the Einstein frame state space that correspond to coordinate singularities in the Jordan frame, thereby emphasizing the importance of physical interpretation of fixed points. This is followed by global considerations, which yield a complete description of the solution space in the Einstein frame. The section ends with situating the global Einstein frame state space in the global Jordan frame state space by means of the variable transformations that link the two approaches, given in Appendix~\ref{app:relations}. This allows us to identify (a) the solutions in the Einstein frame that can be conformally extended in the Jordan frame, and (b) the solutions in the Einstein frame whose evolution completely describes that in the Jordan frame. In Section~\ref{sec:disc} we comment on the relationship between our global Jordan state space approach and other Jordan state space formulations, which are briefly reviewed in Appendix~\ref{app:systems}, where their advantages and disadvantages are discussed. We also give a fairly general discussion of $f(R)$ cosmology, which situates the present models in this more general context.

\label{sec:disc} The global regular system we have presented for the Jordan frame naturally conveys the global properties of the models at hand, as illustrated by Figure~\ref{fig:VacuumJF}. Nevertheless, it is by no means an optimal dynamical system for all aspects one might want to investigate: there are other dynamical systems that have complementary properties. This is already exemplified by the fact that our dynamical system for the Einstein frame, among other things, simplifies the asymptotic analysis of the inflationary attractor solution originating from the $\mathrm{dS}$ fixed point and offers various approximation schemes for the oscillatory regime at late times, thereby complementing other heuristic Jordan frame methods~\cite{mijetal86,LivRev10}. Another useful system is discussed in Appendix~\ref{app:systems} (where its close relationship to the works in~\cite{barher06,car15} is also commented on). It is based on a variable transformation from $(H, R)$ to the variables $(z,q)$, defined by \begin{equation}\label{Defzq} z = \frac{1}{12\alpha H^2}, \qquad q = 1- \frac{R}{6H^2}, \end{equation} and the time variable $N=\ln a$, which leads to the following simple regular system of unconstrained equations: \begin{subequations}\label{zqsys} \begin{align} \frac{dz}{dN} &= 2(1+q)z,\\ \frac{dq}{dN} &= z - \frac32(1-q^2). \end{align} \end{subequations} As discussed in Appendix~\ref{app:systems}, the variable transformation breaks down at $H=0$, which for $z$ and $q$ are located at infinity, i.e., $z$ and $q$ are unbounded. Straightforward compactifications such as a Poincar{\'e} compactification of $z$ and $q$ are inappropriate, since such compactifications result in an erroneous state space topology, which may result in wrong conclusions about the properties of the solutions. For example, such a compactification ruins, or at least complicates, a treatment of the oscillatory regime at late times. \emph{This illustrates that it is necessary to take into account the global topological properties of the physical state space in order to obtain a correct description of the solution space and its properties}, which illustrates a non-local aspect (apart from fixed points reflecting coordinate singularities) concerning the relationship between dynamical systems formulations and (iv): physical solution space interpretation. Nevertheless, the above system has local advantages. The system admits two fixed points, both located on the invariant boundary $z=0$ ($H\rightarrow+\infty$): \begin{subequations}\label{FP_qz} \begin{alignat}{2} \mathrm{R}\!:\quad z &=0, &\qquad q &= 1, \\ \mathrm{dS}\!: \quad z &=0, &\qquad q &= -1. \end{alignat} \end{subequations} The fixed point $\mathrm{R}$ is a hyperbolic source and corresponds to $\mathrm{R}$ in our global system, while $\mathrm{dS}$ is non-hyperbolic with one negative eigenvalue and one zero eigenvalue. % This fixed point corresponds to the fixed point $\mathrm{dS}$ in the global system, obtained after a blow up. A center manifold analysis associated with the zero eigenvalue of $\mathrm{dS}$ yields the following approximation for the inflationary attractor solution (see Appendix~\ref{app:systems}): \begin{equation}\label{qzcenter} q(z) = -1 + \frac{z}{3}\left[1 + \frac{z}{3} - \frac{z^3}{3^5} + \dots\right]. \end{equation} Note that obtaining an approximation for the inflationary attractor solution that comes from $q=-1$ when $H\rightarrow \infty$ is considerably easier in these variables than for the $T$, $\theta$ variables, and the expansion is given as $q(H^{-2})$ since $z \propto H^{-2}$, which might be regarded as preferable. The above brings the inflationary attractor solution into focus. Usually the inflationary regime is understood in terms of slow-roll approximations. In the Jordan frame this approximation can be found in~\cite{mijetal86,LivRev10} and reads $\dot{H} = -1/36\alpha$. Since $\dot{H} = -(1+q)H^2$, this leads to $1+q= 1/36\alpha H^2$, which gives $q = -1 + z/3$, i.e., it just gives the leading order term in the center manifold expansion~\eqref{qzcenter} for the inflationary attractor solution in the variables $z,q$. A comparison with the global variables shows that to leading order $z = (\theta/2)^2$. The above illustrates that not only are the $z,q$ variables a useful complement to the global variables $T,\theta$, since they more straightforwardly give approximations for the inflationary regime, but they are also intimately linked to the Hubble slow-roll approach. Next we consider the Einstein frame and the usual slow-roll approximation. In this setting the slow-roll approximation is obtained by inserting $\tilde{H}=\kappa\sqrt{V(\phi)/3}$ into \begin{equation}\label{dotphiHphi} \kappa\frac{d\phi}{d\tilde{t}} = - 2\frac{\partial \tilde{H}}{\partial \phi}, \end{equation} which for the present scalar field potential gives \begin{equation}\label{SR} \kappa\frac{d\phi}{d\tilde{t}} \approx -\sqrt{\frac{2}{3}} M e^{-\sqrt{\frac{2}{3}}\kappa\phi} . \end{equation} Expressed in terms of the variables $\Sigma_\phi$ and $\tilde{X}$, this results in \begin{equation}\label{slow-roll} \Sigma_\phi \approx -\frac{1}{3} \left[\left(\frac{\tilde{T}}{1-\tilde{T}}\right) - 2\tilde{X}\right]. \end{equation} In the neighborhood of $\mathrm{dS}$, represented by $\tilde{\theta}=\tilde{T}=0$, this yields \begin{equation} \tilde{\theta} \approx -3\left(\tilde{T}-\frac{2}{3}\right), \end{equation} which is the tangency condition for the center submanifold of $\mathrm{dS}$, given by the leading order expression in~\eqref{ThetaExpu}. The slow-roll approximation is therefore just an approximation for the center manifold in the vicinity of $\mathrm{dS}$ in our Einstein frame state space formulation. In this context it should be pointed out that we can of course use variable relationships, given in Eq.~\eqref{JorEinvar} in Appendix~\ref{app:relations}, to translate the various approximations from the Einstein to the Jordan frames and vice versa, and their series expansions can be improved by taking Pad{\'e} approximants, as discussed in e.g.~\cite{alhugg15a}. We end this discussion by emphasizing once more that the main purpose of the presently studied models was to specifically illustrate some general aspects of $f(R)$ cosmology with a simple example, namely (a) the ingredients (i) -- (iv) in the introduction, and (b) some dynamical systems methods with a broad range of applicability, even though the particulars have been tailored to the specific properties of the $f=R + \alpha R^2$ RW models. Although quite special, these models also capture very clearly some central issues in $f(R)$-gravity beyond the above methodological aspects. For example, as stated in the introduction, one way of classifying $f(R)$-gravity models is according to if $F>0$ for all $R$ or not. In the latter case the correspondence between the original $f(R)$ model and its Einstein frame formulation, or its Brans-Dicke ($\omega_\mathrm{BD}=0$) version (see e.g.~\cite{LivRev10,aveetal16} for a description of this correspondence), only holds locally for the range of $R$ where $F>0$. For such models the evolution in the Jordan frame of some solutions are incompletely described in these formulations, i.e., a local formulation correspondence does not entail a global correspondence, as is clearly illustrated in Figure~\ref{fig:Jordan_EinsteinBound}. In this context, note that if $F=0$ was an invariant subset in the Jordan frame, then the $F<0$ part of the Jordan state space would constitute an invariant subset. In this case, one could perhaps argue that the solutions associated with this part of the state space could be discarded on some claimed physical grounds, thus leading to a global physical correspondence between the solution spaces of the different frame formulations. However, as we have shown, $F=0$ is not in general an invariant subset, nor is therefore $F<0$. Thus if one wants to argue that a global physical correspondence exists for the different formulations, one is forced to come up with some arguments for why part of some solutions in the Jordan frame should be discarded (note that the existence of such solutions is ensured by that $F=0$ is not an invariant subset; again, see Figure~\ref{fig:Jordan_EinsteinBound} as an illustrative example). There are, of course, some things that the present models cannot address. In particular this holds for models where the condition $F_{,R}>0$ is broken (leading to e.g. tachyonic instabilities, see e.g.~\cite{LivRev10} and references therein). The change of sign of $F_{,R}$ is particularly problematic from a mathematical point of view since the constraint~\eqref{constJorf} becomes degenerate when $F_{,R}=0$, and, moreover, the causal properties of the field equations change when $F_{,R}$ changes sign (hence the tachyonic instability). There has been some work in $f(R)$ cosmology to extend solutions when the equations are ill-defined, notably~\cite{barher06}. However, we here point out that the existence of $F_{,R}=0$ state space boundaries mathematically resemble sonic shock waves for fluids. It is therefore worthwhile to note that such problems have been dealt with in e.g. the context of spherically symmetric self-similar perfect fluid models~\cite{goletal98a}--\cite{caretal01}, where it was shown how to extend solutions through sonic shock wave surfaces. Incidentally, these models also provide examples where it is useful to cover the state space with several coordinate patches in order to exploit special structures in different parts of the state space, a problem one will inevitably will have to deal with when it comes to most $f(R)$ cosmological models. \subsection*

1607.05715

1607

1607.02629_arXiv.txt

{Most investigations of the X-ray variability of active galactic nuclei (AGN) have been concentrated on the detailed analyses of individual, nearby sources. A relatively small number of studies have treated the ensemble behaviour of the more general AGN population in wider regions of the luminosity-redshift plane.} {We want to determine the ensemble variability properties of a rich AGN sample, called Multi-Epoch XMM Serendipitous AGN Sample (MEXSAS), extracted from the fifth release of the XMM-Newton Serendipitous Source Catalogue (XMMSSC-DR5), with redshift between $\sim 0.1$ and $\sim 5$, and X-ray luminosities in the 0.5-4.5 keV band between $\sim 10^{42}$ erg/s and $\sim 10^{47}$ erg/s.} {We urge caution on the use of the normalised excess variance (NXS), noting that it may lead to underestimate variability if used improperly. We use the structure function (SF), updating our previous analysis for a smaller sample. We propose a correction to the NXS variability estimator, taking account of the light curve duration in the rest frame on the basis of the knowledge of the variability behaviour gained by SF studies.} {We find an ensemble increase of the X-ray variability with the rest-frame time lag $\tau$, given by $\SF\propto\tau^{0.12}$. We confirm an inverse dependence on the X-ray luminosity, approximately as $\SF\propto L_X^{-0.19}$. We analyse the SF in different X-ray bands, finding a dependence of the variability on the frequency as $\SF\propto \nu^{-0.15}$, corresponding to a so-called softer when brighter trend. In turn, this dependence allows us to parametrically correct the variability estimated in observer-frame bands to that in the rest frame, resulting in a moderate ($\lesssim 15\%$) shift upwards (V-correction).} {Ensemble X-ray variability of AGNs is best described by the structure function. An improper use of the normalised excess variance may lead to an underestimate of the intrinsic variability, so that appropriate corrections to the data or the models must be applied to prevent these effects. }

Variability is a distinctive feature shared by all classes of active galactic nuclei (AGN), occurring in all the wavebands and on different timescales from a fraction of a day up to years. In the X-ray band, variability is observed on timescales as short as hours, giving insight into the innermost AGN regions, but also on longer timescales, where variability is seen to increase up to at least a few years \citep[see e.g.][]{mark04,vagn11,shem14}. A large number of studies have investigated the detailed properties of the X-ray variability for many individual AGN, mostly at low redshifts and luminosities \citep[e.g.][]{uttl02,uttl05,pont12}. In cases with sufficient sampling and high signal-to-noise ratios, power spectral density (PSD) analyses have evidenced the typical red-noise character of X-ray variability \citep{gree93,lawr93}. For AGN in wider intervals of redshift and luminosity, including luminous quasars, faint fluxes and sparse sampling usually prevent detailed individual variability studies, nevertheless, average properties of the X-ray variability have been investigated in several ensemble analyses \citep[e.g.][]{alma00,mann02,paol04,mate07,papa08,vagn11}. Different methods are used to estimate the variability of these sources and one of the most popular is the normalised excess variance (NXS), which is defined as the difference between the total variance of the light curve and the mean squared error that is normalised for the average of the $N$ flux measurements squared \citep[e.g.][]{nand97,turn99}; see Sect.\,3. This estimator provides an easy way to quantify the AGN variability even for poorly sampled light curves. However, \citet{alle13} have shown that NXS represents a biased estimator of the intrinsic light curve variance, especially when used for individual, sparsely sampled light curves, which results in overestimates or underestimates of the intrinsic variance that depend on the sampling pattern and the PSD slope below the minimum sampled frequency. Moreover, it has been pointed out that NXS also depends on the length of the monitoring time interval from the red-noise character of the PSD, and decreasing with redshift from the effect of cosmological time dilation \citep[e.g.][] {lawr93,papa08,vagn11}. The structure function (SF) allows one to compute variability as a function of the rest-frame time lag, and is therefore suitable for ensemble analyses. In \citet[][Paper I]{vagn11}, for example, we used multi-epoch observations of an AGN sample extracted from the XMM-Newton serendipitous source catalogue (XMMSSC) to compute the ensemble X-ray SF. In the present paper, we take advantage of the recent releases of XMMSSC \citep{rose16}, and of the Sloan Digital Sky Survey (SDSS) Quasar Catalogue \citep[][P\^{a}ris et al. in preparation]{pari14}, to compute the normalised excess variance and to update the study of the structure function. Moreover, we show that the latter can be also used to correct the time dilation effect present in the estimates of the former. The paper is organised as follows. Section 2 describes the data extracted from the archival catalogues. Section 3 computes the light curve duration effect on the NXS estimates. Section 4 updates the SF computation for the new samples. Finally, in Section 5, we discuss and summarise the results. Throughout the paper, we adopt the cosmology $H_0= 70 {\rm \,km\,s^{-1}\,Mpc^{-1}}$, $\Omega_m=0.3$, and $\Omega_\Lambda=0.7$.

The normalised excess variance is popularly used as a variability estimator. In most cases the method is applied correctly, using monitoring time intervals of fixed duration, for AGN samples at low redshift \citep[e.g. as in][]{pont12}. But this estimator depends on the length of the time interval in the rest frame and is therefore affected also by the cosmological time dilation \citep[e.g.][]{gask81}. The method is sometimes used improperly, choosing non-uniform time intervals, and/or including high redshift sources \citep[e.g.][]{lanz14}, thereby underestimating their variability. A few other examples of this include the work by \citet{la-f14}, which applies NXS to the same data as \citet{pont12} to derive a luminosity distance estimator, but envisages an extension of the study to higher redshift sources, where NXS would underestimate variability. The work by \citet{cart15} applies NXS to the Quest-La Silla variability survey, including high redshift AGNs, whose variability is therefore underestimated. However, their main implication is a trend indicating that high redshift and more variable AGNs tend to have redder colours and this trend would be reinforced taking the NXS underestimate into account. To demonstrate the duration effect for the NXS estimates, we used a sample of AGNs with multi-epoch X-ray observations (MEXSAS) extracted from the fifth release of the XMM-Newton Serendipitous Source Catalogue (XMMSSC-DR5). We have also shown that the effect can be corrected on the basis of the knowledge of the behaviour of variability that is gained from structure function studies; our correcting formula, Eq. 9, can be successfully applied to further NXS-based studies. We have updated the analysis of the ensemble structure function, finding that X-ray variability is well described by a power-law function of the rest-frame time lag, increasing as $\tau^{0.12}$ and extending up to $\sim 2000$ days. We have also shown that X-ray variability is inversely correlated with X-ray luminosity, approximately as $L_X^{-0.19}$. This anti-correlation has been reported, usually at short timescales, by many authors (e.g. \citet{barr86}, \citet{lawr93} for low-$z$ AGNs, \citet{mann02}, \citet{papa08} for higher $z$) with variability approximately proportional to $L_X^{-0.3}$. At longer timescales, the analysis by \citet{mark04}, for local AGNs, indicates $F_{var}\propto L_X^{-0.13}$. One simple interpretation of the anti-correlation is the superposition of several independently flaring subunits \citep[e.g.][]{gree93,nand97,alma00}. We also find a dependence of variability on the emission frequency approximately as $\nu^{-0.15}$. In turn, this dependency is related to the change of the photon index, indicating a softer when brighter spectral variability behaviour, which extends a trend previously found for Seyfert galaxies \citep{sobo09} to AGNs with higher redshifts and luminosities. Because of this dependence, variability in the rest frame differs from that estimated in the observer-frame bands; however the effect can be corrected and we propose a simple correction term called V-correction, resulting in a moderate shift upwards ($\lesssim 15\%$) for the structure function. The same correction, applied in different bins of redshift, can affect the resulting $z$-dependence of variability, suggesting a weak increase with $z$ for the variability at short time lags. We finally remark that the corrections proposed by \citet{alle13} on the NXS should also be taken into account in the case of sparse sampling and for a comparison with physical models.

1607.02629

1607

1607.08568.txt

Slowly rotating magnetic massive stars develop ``dynamical magnetospheres'' (DM's), characterized by trapping of stellar wind outflow in closed magnetic loops, shock heating from collision of the upflow from opposite loop footpoints, and subsequent gravitational infall of radiatively cooled material. In 2D and 3D magnetohydrodynamic (MHD) simulations the interplay among these three components is spatially complex and temporally variable, making it difficult to derive observational signatures and discern their overall scaling trends. Within a simplified, steady-state analysis based on overall conservation principles, we present here an ``{\em analytic dynamical magnetosphere}'' (ADM) model that provides explicit formulae for density, temperature and flow speed in each of these three components -- wind outflow, hot post-shock gas, and cooled inflow -- as a function of colatitude and radius within the closed (presumed dipole) field lines of the magnetosphere. We compare these scalings with time-averaged results from MHD simulations, and provide initial examples of application of this ADM model for deriving two key observational diagnostics, namely hydrogen H-$\alpha$ emission line profiles from the cooled infall, and X-ray emission from the hot post-shock gas. We conclude with a discussion of key issues and advantages in applying this ADM formalism toward derivation of a broader set of observational diagnostics and scaling trends for massive stars with such dynamical magnetospheres.

Hot luminous, massive stars of spectral type O and B have dense, high-speed, radiatively driven stellar winds \citep{Castor75}. In the subset ($\sim$10\%) of massive stars with strong ($> 1$kG), globally ordered (often significantly dipolar) magnetic fields \citep{Petit13}, the trapping of the wind outflow by closed magnetic loops leads to the formation of a circumstellar {\em magnetosphere}. For stars with moderate to rapid rotation -- such that the Keplerian corotation radius $R_K$ lies within the Alfv\'{e}n radius $R_A$ that characterizes the maximum height of closed loops --, the rotational support leads to formation of a ``{\em centrifugal magnetosphere}'' (CM), wherein the trapped wind material accumulates into a relatively dense, stable and long-lived rigidly rotating disk \citep{Townsend05}. In contrast, for magnetic massive stars with slow rotation, and thus $R_K > R_A$, this trapped material falls back to the star on a dynamical (freefall) timescale, representing then a ``{\em dynamical magnetosphere}'' (DM) \citep{Sundqvist12c, Petit13}. Because of the rotational spindown associated with angular momentum loss through the relatively strong, magnetized wind upflow from open field regions \citep{Uddoula09}, magnetic O-type stars are typically\footnote{The one exception is Plasket's star, for which the magnetic star likely has been spun up by mass accumulation from its binary companion \citep{Grunhut13}.} slow rotators, and so harbor DM's. Among the magnetic B-stars, a significant fraction (about half) are also rotating slowly enough to have DM's \citep{Petit13}. In such DM's the trapped wind upflow from opposite footpoints of closed loops collides near the loop apex, forming shocks that heat the gas to X-ray emitting temperatures; as this gas radiatively cools, it falls back to the star as a gravitational downflow. 2D and 3D magnetohydrodynamic (MHD) simulations of such DM's \citep{Uddoula02,Uddoula13} show a complex and variable interplay among all three components, and this makes it difficult to derive observational signatures and discern their overall scaling trends. Applications of these MHD models have thus been limited to a few selected O-stars, using simplified radiative transfer methods to derive synthetic spectra for H$\alpha$ emission lines \citep{Sundqvist12c, Grunhut12, Uddoula13, Wade15} and ultra-violet wind resonance lines \citep{Marcolino13, Naze15}. These initial studies have provided strong support of the general DM concept; however, the complexity of the time-dependent 3D structure, together with computational expense of the simulations, prohibits more systematic and detailed computations of synthetic observables across the electromagnetic spectrum, as well as application to larger samples of magnetic massive stars. Similar arguments apply to X-ray spectral synthesis. Detailed MHD simulations have been used to analyze the high spectral resolution X-ray observations available for a few selected OB stars \citep{Petit15,Naze15}. But for the analysis of the much larger number of stars with low-resolution X-ray data, an analytic model can capture key observable properties and trends. Recently, \citet[][hereafter paper I]{Uddoula14} carried out an extensive MHD simulation study of radiative cooling of the hot post-shock gas in DM's, with a focus on deriving the resulting X-ray emission as a function of the density-dependent cooling efficiency. When interpreted in terms of a simplified ``X-ray analytic dynamical magnetosphere" (XADM) analysis, this led to predicted scaling laws for variation of X-ray luminosity with the wind mass loss rate. Subsequent application by \citet{Naze14} toward interpreting observationally inferred X-ray luminosities in a large sample of magnetic massive stars showed that, with some fixed factor adjustment to account for the limited ``duty cycle'' for X-ray production during the complex variations seen in MHD simulations, this XADM scaling matched quite well the observed trends for the subset of magnetic massive stars with slow enough rotation to have DM's. The analysis here builds on this success to develop a more general ``analytic dynamical magnetosphere'' (ADM) formalism that now provides explicit formulae for the spatial variation of density and flow speed (as well as temperature for the hot gas) in all three components of the closed loop magnetosphere: wind upflow, hot post-shock gas, and cooled downflow (section \ref{sec:adm}). The overall ADM analysis here is guided and tested by comparison with time-averaged results from full MHD simulations of DM's (section \ref{sec:MHD}), using the Alfv\'en radius $R_A$ in the MHD simulation to define a maximum closure radius $R_c$ of a dipole loop in the ADM model. The inclusion of the cooled downflow now allows for modeling of optical emission lines like H-$\alpha$ (section \ref{sec:Balmer}). Moreover, the description of the spatial distribution of both the hot and cool components allows an extension of the XADM analysis of paper I (section \ref{sec:xrays}), for example by accounting for possible bound-free absorption of emitted X-rays by the cool wind and downflow \citep[see section \ref{sec:xray-abs} and][]{Petit15}. We conclude (section \ref{sec:conclusions}) with a brief summary of key issues and advantages in applying this ADM formalism toward deriving a broader range of spectral diagnostics. Appendices A and B show how the ADM model can be used to derive both the X-ray differential emission measure (DEM), as well as a shock-temperature distribution $p(T_s)$. Appendix C gives background on the ADM scalings for H$\alpha$ emission. \begin{figure} \begin{center} \vfill \includegraphics[scale=0.48]{ADM-fig1.pdf} \caption{ Illustration of three components of material flow along a closed dipole loop line that intersects the stellar radius $\Rstar$ at a colatitude $\theta_\ast = \arccos \mustar$, and extends up to a maximum radius $r_m = \Rstar/ (1-\mustar^2)$ at the equatorial loop apex. Wind upflow (black arrows) from a stellar surface footpoint meets wind material from the opposite footpoint at the loop apex. This results in a pair of reverse shocks with hot post-shock gas (red) extending back from the loop apex to shock fronts at co-latitudes $ \pi - \theta_s$ and $\theta_s (\equiv \arccos \mu_s )$, and radius $r_s = r_m (1-\mu_s^2)$. Cooled material at the loop apex (blue wedge) is pulled back by the stellar gravity, forming channels of cooled downflow (blue dashed arrows) back toward the star. } \label{fig:dipgeom} \end{center} \end{figure}

\label{sec:conclusions} The ADM analysis here provides readily computable formulae for the basic hydrodynamic quantities -- density, temperature, and velocity -- for each of the three components of a wind-fed dynamical magnetosphere -- wind upflow, hot post-shock gas, and cooled downflow. Comparison with time-averaged values derived from detailed MHD simulations show, with some caveats, quite good general agreement. As such, this ADM formalism can provide a conceptually and computationally much simpler basis for synthesizing observational diagnostics, and for deriving broad scaling relations for how these depend on stellar, wind, and magnetic parameters. For X-ray spectral bands, \S \ref{sec:xray-adm-vs-mhd} compares directly the X-ray emission in ADM vs.\ MHD models, while \S \ref{sec:xray-abs} discusses how observed X-ray spectra could be affected by absorption from the cool components. Appendices A and B present a general scaling analysis of how the ADM model can be used to derive both the differential emission measure, as well as a shock-temperature distribution $p(T_s)$. This augments the XADM analysis of paper I, and so builds on the promising agreement of the derived scaling laws with observations \citep{Uddoula14,Naze14}. To illustrate ADM spectral synthesis in optical emission lines, we exploit the relative simplicity of the H$\alpha$ line formation process in O-type stars. \S~\ref{sec:Balmer} explicitly demonstrates the potential diagnostic power of the ADM model by a first successful application to observations of the rotational phase variation of H$\alpha$ emission from the O-star HD191612. The remarkably good agreement provides constraints on key physical parameters like magnetic geometry and overall mass loss rate $\dot{M}_{\rm B=0}$, thus illustrating the utility of the ADM formalism even for modeling individual stars with DM's. The analysis in Appendix C provides also general scaling relations with stellar, wind, and magnetospheric parameters. But the simple, steady-state nature of the ADM model paves the way for future applications that require more elaborate NLTE radiative transfer. For example, in magnetic O-stars optical helium lines like HeII 4686${\rm \AA}$ show clear signatures of being formed in a DM \citep[][]{Grunhut12, Wade15}, and recent observations of magnetic massive stars in the infra-red \citep[e.g.,][]{Oksala15} suggest a strong influence of the magnetosphere also for key diagnostics in that waveband. Moreover, the physical explanation of the so-called Of?p morphological phenomena \citep{Walborn72} of magnetic O-stars is very likely related to a complex formation scenario of nitrogen and carbon spectral lines in a DM. In summary, much as complex computer codes like {\sc cmfgen} \citep{Hillier98} or {\sc fastwind} \citep{Puls05} nowadays are routinely applied for spectroscopic analyses of non-magnetic hot stars with winds, we envision that the ADM model presented here lays the groundwork for development of NLTE radiative transfer tools for detailed spectroscopic analysis of magnetic massive stars.

1607.08568

1607

1607.05186_arXiv.txt

Very wide binaries ($> 500$ AU) are subject to numerous encounters with flying-by stars in the Galactic field and can be perturbated into highly eccentric orbits ($e \sim 0.99$). For such systems tidal interactions at close pericenter passages can lead to orbit circularization and possibly mass transfer, consequently producing X-Ray binaries without the need for common envelope. We test this scenario for the case of Black Hole Low-Mass X-Ray Binaries (BH LMXBs) by performing a population synthesis from primordial binaries with numerical treatment of random stellar encounters. We test various models for the threshold pericenter distance under which tidal forces cause circularization. We estimate that fly-by interactions can produce a current population of $\sim60$--$220$ BH LMXBs in the Galactic field. The results are sensitive to the assumption on tidal circularization efficiency and zero to very small BH natal kicks of a few km/s are required. We show that the most likely donors are low-mass stars (< 1 $\rm M_{\odot}$; at the onset of mass transfer) as observed in the population of known sources ($\sim 20$). However, the low number of systems formed along this route is in tension with most recent observational estimate of the number of dormant BH LMXBs in the Galaxy $10^4$--$10^8$ \citep{Tetarenko2016}. If indeed the numbers are so high, alternative formation channels of BHs with low-mass donors need to be identified.

There are currently 19 Galactic X-ray binary systems in which black holes (BHs) have been confirmed dynamically, with a few more extragalactic sources observed \citep{Remillard2006, Casares2014, Wiktorowicz2014, Tetarenko2016B}. They are all close binaries (with the majority of orbital periods $P_{\rm orb} < 1$ day), for which a mass transfer occurs and an accretion disc is formed around the black hole and is responsible for the X-ray activity (\citealt{Shakura1973}; \citealt{Lasota2015}). 16 of these systems possess companion stars of relatively low mass ($<2 \,\rm M_{\odot}$), with the distribution peaking at around $0.6\, M_{\odot}$ and spectral types ranging from A2V to M1V -- they are classified as low-mass X-ray binaries (LMXBs). All known BH LMXBs are transient X-ray sources, exhibiting occasional outbursts with their brightness increasing by 3-5 orders of magnitude, which are attributed to disc instability \citep[e.g.][]{Lasota2001}. The BH LMXBs are most commonly thought to originate from primordial binaries of a black hole progenitor \citep[$M_{\rm a} > 20-25 \, \rm M_{\odot}$, e.g.][]{Fryer2012} and a much less massive main sequence (MS) secondary ($M_{\rm b} < 2 \, \rm M_{\odot}$) although evolution starting with intermediate mass secondaries was also proposed \citep{Justham2006, Chen2006}. As the primary expands during its evolution it is expected to fill its Roche lobe and launch a dynamically unstable mass transfer, initiating the common envelope (CE) phase \citep{Paczynski1976}. As a result of this short-lived evolutionary stage the separation between the components decreases significantly, which ultimately leads to a later mass transfer from low-mass star to a black hole and prolonged X-ray activity. The formation of a black hole from a Zero-Age Main-Sequence (ZAMS) star takes less than 10 Myr. In such a time the low-mass secondary is still at the beginning of the MS \citep[or even still during pre-MS contraction;][]{Ivanova2006} and is expected to remain a MS star for the next $2.5$--$10$ Gyr. This now considered standard scenario was first suggested nearly 30 years ago to explain the origin of the A0620-00 black hole X-ray binary \citep{deKool1987}. Since then, however, multiple studies have pointed out flaws of this conception in the particular case of BH LMXBs, suggesting that it is difficult for a binary of a massive giant and a much less massive companion to eject the CE at the expense of its orbital energy \citep[e.g.][]{Podsiadlowski2003}. Thus, unless abnormally high values for the CE ejection efficiency are adopted \citep{Yungelson2008}, the system is most likely to merge. In an attempt to justify such increased efficiency several modified CE models were proposed (see \citealt{Li2015} and references therein). Population synthesis studies have recently shown, however, that CE models in their current state fail to reproduce the distribution of donor mass of the observed BH LMXBs (\citealt{Wiktorowicz2014}; \citealt{Wang2016}), implying the need for either a more significant adjustment of the CE modeling or an alternative formation scenario to be considered. Recently \citealt{Michaely2016} (hereafter MP2016) suggested that LMXBs (including those with black holes) could originate from very wide binaries ($a > 1000\,\rm AU$) that undergo a series of subsequent fly-by interactions with stars of the Galactic field. The authors predict that some of these wide systems should be excited into orbits of very high eccentricity, eventually circularized by tidal effects at pericenter. This essentially transforms a wide binary into a compact one, playing a similar role to the CE phase in the standard scenario. Dynamical formation of LMXBs is definitely not a new concept when it comes to dense stellar environments, such as globular clusters, where gravitational interactions and tidal captures take place most often (\citealt{Clark1975}; \citealt{Fabian1975}). In fact, it appears that formation channel of X-ray binaries through dynamical processes is very efficient, as $\sim 10\%$ of all LMXBs is being found in globular clusters, which contain only about $\sim 0.1 \%$ of all stars in the Galaxy \citep{Irwin2005} (note that these are all X-ray binaries with a neutron star as the accretor as all known BH LMXBs reside outside globular clusters). Similar processes should be effective also in the most dense, central regions of galaxies. This was confirmed by \citet{Voss2007} who carried out Monte Carlo simulations of a binary population in the bulge of M31, applying the \textsc{fewbody} code \citep{Fregeau2004} to numerically simulate random dynamical interactions with passing by stars. A similar study, however, has never been done for the Galactic disc, where all of the known BH LMXBs reside \citep{Li2015}. Although the disc is typically considered to not be dense enough for stellar encounters to play an important role, dynamic interactions do become significant for wide binaries with naturally larger cross-sections (e.g. \citealt{Bahcall1985}). It is even predicted that very wide binaries should be the primary source of stellar collisions in the Milky Way \citep{Kaib2014}, thanks to similar excitations to highly eccentric orbits as considered by MP2016 in their scenario for LMXBs formation. Here we perform a population synthesis study on the dynamical formation of BH LMXBs from primordial binaries in the Galactic disc. We focus only on binaries with a black hole as their primary and a MS star as their secondary, as the origins of these systems remain most mysterious \citep[MS companions are also most frequent among observed systems, as $\sim75\%$ of donors are low-mass dwarf stars,][]{CorralSantana2016}. The paper is organized as follows: in section 2 we describe the MP2016 scenario, as well as introduce our modeling of dynamic interactions and tidal circularization. In section 3 we describe our population synthesis approach and utilized computer codes. In section 4 we present and discuss our results. We conclude in section 5.

\begin{enumerate} \item \textbf{The number of BH LMXBs:} we find that dynamical formation of LMXBs from primordial binaries in the Galactic field can lead to a population of $\sim$ 60--220 BH LMXBs, depending on the model of tidal circularization. This number could possibly be higher by a factor of about 4.5 if the distribution of initial mass ratios was $f(q) \sim q^{-0.5}$ \citep{Duchene2013}, instead of the flat distribution we adopted. In the same time though, we do not account for the finite duration of the LMXB phase and we simplistically assume initial mass function for the mass function of flying-by field stars. Both those assumption lead to an overestimation in the number of BH LMXB candidates, possibly be a factor of 5 and 2.5 respectively (see Sec.~\ref{sec:lifetimes} and~\ref{sec:mass_func_flybys}). Our resulting population is most likely a small fraction of the expected total number of BH LMXBs in the Milky Way, as the estimates based on observational surveys typically span over the range of $10^2$--$10^4$ systems (e.g. \citealt{Romani1998}; \citealt{CorralSantana2016}), whereas the recent discovery of the first candidate for quiescent BH LMXB outside of a globular cluster by \citet{Tetarenko2016} indicates, that quiescence BH LMXB may be even more abundant in the Galaxy than previously thought, rising the estimate of the total number of BH LMXBs up to $10^4$--$10^8$. Our results are based on a simple treatment of tidal circularization during sufficiently close periastron passages, where we assume that mass transfer will be launched in all tidally circularized systems, leading to a LMXB formation in each case, which, in reality, might not always be the case \citep{Ray1987}. Thus, our estimates serve more as upper limits and we consider it unlikely that a more detailed study could obtain a much bigger population of $\sim 10^4$ dynamically formed BH LMXBs. Our results rather indicate that the dynamical formation channel is only responsible for a small fraction of all the Galactic sources -- unless the population of wide BH-MS binaries is much larger than we considered. Possible additional sources of such systems include ultra-wide primordial binaries with massive BH progenitors (with orbital periods beyond the limit $\rm log(P/day) = 5.5$ we assumed), which could increase the number of dynamically formed BH LMXBs by a factor of about 3.5--4 (see Sec.~\ref{sec:ultra_wide_section}), as well as the cluster-dispersal scenario in which a massive object (such as a stellar BH) tidally captures a low-mass star on a wide orbit following a dispersal of their host cluster \citep{Perets2012}. Another possibility for dynamical formation of LMXBs are channels involving triple star systems \citep[eg.][]{Perets2012b,Naoz2016}. \item \textbf{Important factors :} We find that wide ($ \gtrsim 200 \rm \; AU$), fast moving BH-MS binaries ($40-110 \; \rm km \; s^{-1}$ relative to local Galactic velocity) are most likely candidates for dynamical formation of BH LMXBs. For such systems the frequency of stellar encounters is sufficient (on average several thousand interactions during a few Gyr evolution), so that it is possible for their orbits to be perturbated into highly eccentric states and then circularized by tidal forces at close pericenter passages. In principle, the dynamical scenario would most efficient for systems with separations between $2000$ and $6000$, although the existence of BH-MS binaries this wide is uncertain (see Sec.\ref{sec:ultra_wide_section}). Note that high velocities of our systems are solely associated with dispersion of the Galactic velocity distribution, as the natal kicks of BHs in the surviving binaries in our simulations were either very small (a few km/s) or negligible. In fact, we find that the dynamical scenario quickly becomes ineffective with increasing magnitude of BH natal kicks as only about 10\% of our BH LMXB candidates could survive BH formation with kicks drawn from an Maxwellian distribution with $\sigma_{NK} = 10$ km/s (see Sec.~\ref{sec:bh_natal_kicks}). Even though higher kick velocities could potentially produce fast-moving systems with respect to the surrounding stars, giving chance for more stellar encounters and working in favor of the dynamical scenario, we consider this effect to be of secondary importance. We also show that the effectiveness of dynamical formation channel is largely dependent on the effectiveness of tidal circularization, which is very much a matter of debate. A detailed treatment of tidal forces would narrow the constraints on the number of formed BH LMXBs. We find that a typical evolutionary path leading towards a highly eccentric binary is dominated by only a handful ($\sim$ 10) of truly significant interactions (i.e. orbital energy changing by more than 10 \%), even though the total number of encounters is usually of the order of several thousands. These few key interactions are, on average, with exceptionally massive fly-bys (often $>\rm 10 \; M_{\odot}$). \item \textbf{Companion mass function:} we find that the distribution of donor masses in dynamically formed BH LMXB systems is very sensitive to changes of the condition for tidal circularization at close pericenter passage, which is still a poorly understood and difficult to model process. Thus, a more detailed treatment of tidal forces is required to conclude about the exact shape of the CMF in dynamically formed BH LMXB population. We do show, however, that in general the most likely companions are low-mass stars of $\rm M < 1 \rm \; M_{\odot}$ with the distribution peaking at around $\rm \sim 0.9 \; M_{\odot}$ in all our models based on the Roche Lobe filling factor (see table~\ref{tab:first_table} and figure~\ref{fig:mass_distribution}). Since we do not simulate the mass transfer itself, in reality the donor masses in dynamically formed BH LMXBs should be even smaller. This result is in agreement with the observed population of BH LMXBs for which donor masses span mostly in range 0.1--1.0 $\rm M_{\odot}$, peaking at around $\rm \sim 0.6 \; M_{\odot}$ (see \citet{Wiktorowicz2014} and Fig.~\ref{fig:mass_distribution}). \end{enumerate}

1607.05186

1607

1607.06841_arXiv.txt

{HD~208472 is among the most active RS~CVn binaries with cool starspots. Decade-long photometry has shown that the spots seem to change their longitudinal appearance with a period of about six years, coherent with brightness variations.} {Our aim is to spatially resolve the stellar surface of \object{HD~208472} and relate the photometric results to the true longitudinal and latitudinal spot appearance. Furthermore, we investigate the surface differential rotation pattern of the star.} {We employed three years of high-resolution spectroscopic data with a high signal-to-noise ratio (S/N)\ from the STELLA robotic observatory and determined new and more precise stellar physical parameters. Precalculated synthetic spectra were fit to each of these spectra, and we provide new spot-corrected orbital elements. A sample of 34 absorption lines per spectrum was used to calculate mean line profiles with a S/N of several hundred. A total of 13 temperature Doppler images were reconstructed from these line profiles with the inversion code $iMap$. Differential rotation was investigated by cross-correlating successive Doppler images in each observing season.} {Spots on HD~208472 are distributed preferably at high latitudes and less frequently around mid-to-low latitudes. No polar-cap like structure is seen at any epoch. We observed a flip-flop event between 2009 and 2010, manifested as a flip of the spot activity from phase 0.0 to phase 0.5, while the overall brightness of the star continued to increase and reached an all-time maximum in 2014. Cross-correlation of successive Doppler images suggests a solar-like differential rotation that is $\approx$ 15 times weaker than that of the Sun.} {}

\label{S1} Well-sampled photometric monitoring of spotted cool stars allows tracing photospheric brightness variations that are caused by rotational modulation. Long-term photometric data moreover allow investigating spot activity cycles and relate inherent changes of the surface spot distribution to photometric period variations, both of which are an indicator of the underlying surface differential rotation. Phenomena such as active longitudes \citep{henry95,jetsu96} and flip-flops of active longitudes \citep{jetsu91,berd_tuo98} were revealed by such long-term photometry, but they are far from being fully understood. \citet{holz03a,holz03b} demonstrated on theoretical grounds that the tidal effects of a companion star in a binary system may alter the surface distribution of spots by breaking the axial symmetry, which could lead to preferred longitudes. Magnetic dynamo simulations \citep{fluri04,moss05,elstner05} of the flip-flop phenomenon basically concluded that the axisymmetric and the non-axisymmetric (magnetic) field components must coexist to produce a flip-flop in the first place. These studies highlighted that meridional circulation and differential rotation are a prerequisite for this type of dynamo but do not yet allow predicting the relation between differential rotation and flip-flop period, for example. Recently, \citet{rot15a, rot15b} showed that there might be ambiguity in interpreting light curves of RS CVn binaries in terms of persistent active longitudes and ellipsoidal distortion. In this case, some part of the light curve might be caused by binarity effects and not by cool spots. For the characterization of differential surface rotation, time-series Doppler imaging became the most powerful method (see, e.g., \citealt{donati97,strass09,kunstler15,kov16}). The differential rotation signature is either based on the longitudinal cross-correlation of images from successive rotation cycles or on the addition of the image shear as an additional free parameter in the line-profile inversion itself. Neither of these variants is free of problems, see \citet{dun08} for a comparison, but both were successfully applied to a few single and binary stars in different evolutionary stages (e.g., \citealt{don03,dun08,kov12}). For main-sequence stars, only solar-type differential rotation (equator rotates faster than poles) was found so far, while anti-solar differential rotation (poles rotate faster than equator) was found for a few fast-rotating giants \citep{weber05}. \citet{kit_rud04} have shown that fast meridional circulation in a very deep convection zone with magnetically induced thermal inhomogeneities would lead to anti-solar differential rotation. However, at least in some cases, different authors found contradictory results for one and the same target, for example, for the K sub-giant of the RS~CVn binary \object{HR~1099}, for which \citet{vogt99} analyzed high-resolution unpolarized spectroscopic data and found weak anti-solar differential rotation, while \citet{petit04} used high-resolution polarized spectra and found solar-type differential rotation. It is mostly assumed that the spots follow the rotation of the surface, while differential rotation is investigated by photometric spot modeling and/or time-series Doppler imaging. However, \citet{korh11} suggested that large spot structures could reflect geometric properties of the large-scale dynamo and not the surface rotation. Still, the spots are the tracers we rely on. \object{HD~208472} (\object{V2075 Cyg}, \object{HIP 108198}, $V$=7\fm4, $P_{\rm phot}$=22.4\,d) is a conspicuous target that seems to show active spot longitudes and an activity cycle with a period of about six years \citep{ozd10} (hereafter paper~I). In 1991, W.~Bidelman discovered its strong Ca\,{\sc ii} H and K emission, and a number of follow-up studies revealed and confirmed that the system is a single-lined spectroscopic binary with a G8III primary component on a nearly circular orbit that exhibits rotationally modulated light variations with variable light-curve amplitude \citep{henry95,fekel99,strass99,koen_eye02}. It was also suggested that the system may be a good candidate for Doppler imaging \citep{henry95}. First such attempts for Doppler imaging were made by \citet{weber01}, \citet{weber04} and \citet{weber05}, whose initial study was based on 70 consecutive nights of spectroscopic observations with the NSO McMath-Pierce telescope between 1996 and 1997. They found that spots at that time were concentrated at low and intermediate latitudes without any structures close to the one visible pole of the star. They applied the sheared-image method and found a weak anti-solar differential rotation of $\alpha=-0.04\pm0.02,$ which, however, was not considered a conclusive result. \citet{erdem09} estimated an $\alpha$ parameter of the same amount, but relative to the orbital period, through photometric spot modeling based on $BVR_{c}I_{c}$ photometry in 2006. More comprehensive photometry was presented in our paper~I, where we analyzed 17 years of data from Automatic Photoelectric Telescopes (APT). A spot-modeling and photometric-period analysis led us to a differential rotation coefficient of $\alpha=0.004\pm0.010$, which practically indicated a non-detection of differential rotation. Furthermore, we found that the longitudinal position of spots varied coherently with the mean brightness and seemed to exhibit a 6.28\,yr period. This cyclic variation was interpreted as a stellar analog of the solar 11-year sunspot cycle. In the current study, we present new high-resolution time-series spectroscopy of the system from 2009--2011 together with contemporaneous $V$-band photometry. We redetermine precise atmospheric parameters, obtain an improved spectroscopic orbit ,and investigate the evolutionary status of the primary star. The unique time-series data from STELLA enables us to obtain Doppler images of the star for many successive rotational cycles, and for three consecutive years. We first summarize the instrumental setup and the data collection and reduction procedures in Sect.~\ref{S2}. In Sect.~\ref{S3} we derive stellar atmospheric parameters, spectroscopic orbital elements and physical properties of the system. Section~\ref{S4} comprises the Doppler imaging of the primary star, including data preparation and a brief description of the $iMap$ inversion code. We present a total of 13 surface maps and their analysis in Sect.~\ref{S5}. In the final section, we summarize and discuss our results.

Three years of high-resolution time-series STELLA spectra enabled us to not only refine precise spectroscopic orbital elements and atmospheric parameters of HD~208472, but reconstruct 13 separate surface images from three observing seasons through multiple-line inversions. Surface reconstructions were made by tracing the distortions in the absorption line profiles with time, which is the main principle of our time-series approach of Doppler imaging. The line-profile distortions are also the cause of systematic shifts in radial velocities, known as spot jitter in case of chromospherically active stars. After proper removal, it decreased the RV residuals for our new spectroscopic orbital solution by a factor of four from 380~\ms\ to 88~\ms. Fourier analysis of the RV residuals indicated that the dominant periods are always very close to the orbital period and its harmonics, which is the typical signature of spot jitter. HD~208472 exhibits active longitudes on adjacent hemispheres and occasional flip-flops between them that overlap the continuous drift in longitude with respect to the orbital reference frame. Similar active longitudes and flip-flops were observed in several RS~CVn binaries \citep{jetsu96,berd_tuo98,kunstler15}, FK~Com-type stars \citep{jetsu91,jetsu99}, and young solar-like stars \citep{berd05}. There is some evidence that flip-flops occur more or less regularly with timescales of a few years up to a decade, sometimes referred to as flip-flop cycle (see the references above), but the evidence remains inconclusive. To detect these cycle lengths, we still need long-term photometry with phase-resolved sampling or, even better, modern time-series Doppler imaging. A 6.28 yr flip-flop cycle was discovered for HD~208472 in our paper~I, where we predicted the begin of a new cycle for 2009/2010. Our Doppler imagery in the present paper caught this event. We witnessed the change of the main activity zone from $\phi = 0.00$ to $\phi = 0.50$ between the 2009 and 2010 observing seasons over a time range of approximately one year. Thus, the location of the main activity zone corresponds to the sub-stellar point and its antipode, as has been suggested by \citet{olah06}. Theoretically, coexistence of axisymmetric and non-axisymmetric dynamo modes may produce active longitudes and also give rise to the flip-flop phenomenon \citep{fluri04,moss05,elstner05}. However, why these are aligned and influenced by the secondary star and why the brightness continues to increase over several activity cycle is yet to be understood. From cross-correlating successive Doppler maps, we found a surface shear of $\alpha$ = +0.015$\pm$0.003 for a $\sin^2$-latitude differential rotation law and $\alpha$ = +0.012$\pm$0.008 for a $\sin^2 + \sin^4$-latitude differential rotation law. This differential rotation is in solar-like direction, meaning that poles rotate slower than the equator, but $\approx$ 15 times weaker than on the Sun. The average value is close to the $\alpha$ value of \object{XX~Tri} \citep{kunstler15}, which is a similar RS~CVn system in terms of physical properties and spectroscopic orbit, but maybe$\text{ }$\text{about twice as} old as HD~208472. Comparing other stars whose $\alpha$ values were determined from Doppler imaging, our target star belongs to a group of weak solar-type differential rotators (\object{AB~Dor}, \citealt{donati97}; \object{$\zeta$~And}, \citealt{kov12}; \object{IL~Hya}, \citealt{kov14}). On the other hand, anti-solar differential rotation was found (or at least claimed) in some other active stars, such as \object{$\sigma$~Gem} \citep{kov15}, \object{UZ~Lib} \citep{vida07}, \object{HU~Vir} \citep{str94,gohar}, and \object{HD~31993} \citep{strass03}. Tidal force in a binary system is considered as a possible explanation for the existence of two different directions of differential rotation, depending on the orbital and physical properties of the related system. \citet{holz_02} showed that the tidal force in an RS~CVn binary may lead to preferred longitudes. It may also affect the angular momentum distribution in the convective envelope of the giant component and alter its differential rotation pattern \citep{kov12}. Another effect of the tidal force on differential rotation might be the suppression of the strength of the differential rotation through tidal locking \citep{col_cam07}. The RS~CVn binaries mentioned above have $\alpha$ values of just a few of per cent (solar or anti-solar). In comparison, the single K2 giant \object{HD~31993} \citep{strass03}, where no tidal interaction is present, has a dramatically higher shear value of $\alpha = -0.125$. Another comparison can be made with the single K1 giant \object{DP~CVn} \citep{kov13}, which has a surface shear value of $\alpha = -0.035$, which is a stronger surface shear in magnitude than \object{HD~208472}. This shows that fast-rotating single giants with $\alpha$ values even comparable to the solar value have been reported, while no such RS CVn stars have been found. This supports the picture that differential rotation is suppressed by tidal forces as suggested by \citet{schar82}, for example.

1607.06841

1607

1607.08895_arXiv.txt

We use published data on the power and production efficiency of jets in blazars with double radio lobes in order to compare results obtained using different methods. In order to eliminate selection effects, we use cross-matched sub-samples containing only luminous blazars. We compare the three main existing methods, namely those based on the emission of radio lobes, on spectral fitting, and on radio core shift. We find the average jet power obtained for identical samples with the radio-lobe method is $\sim$10 times lower than that from the spectral fitting. In turn, the power from spectral fitting is compatible with that from core-shift modelling for plausible parameters of the latter. We also consider a phenomenological estimator based on the \g-ray luminosity. We examine uncertainties of those methods and discuss two alternative hypotheses. In one, the blazar-fit and core-shift methods are assumed to be correct, and the lower power from radio lobes is caused by intermittency of accretion. Long periods of quiescence cause the energy in the radio lobes, accumulated over the lifetime of the blazar, to be much less than that estimated based on the present luminous state. In addition, the power calculated using the radio lobes can be underestimated for intrinsically compact jets, in which the radio core flux can be over-subtracted. In our second hypothesis, the radio-lobe method is assumed to be correct, and the blazar-fit and core-shift powers are reduced due to the presence of $\sim$15 pairs per proton and a larger magnetization than usually assumed, respectively.

Starting with the seminal paper by \citet{Rawlings1991}, it has become common to use energetics and lifetimes of extended double radio sources to calculate jet powers in radio galaxies and quasars. Such studies indicated that in radio-loudest objects the jet power, $P_{\rm j}$, is comparable or even exceeds the accretion power, $\dot M c^2$, where $\dot M$ is the accretion rate (see, e.g., \citealt{Punsly2007, Fernandes2011, Sikora2013}). Problems with launching such powerful jets by standard accretion discs stimulated studies on the so-called magnetically arrested discs (MAD; \citealt{Narayan2003}). In scenarios involving the MAD, the jet is powered by a fast spinning black hole (BH) immersed in a strong magnetic field supported by the ram pressure of the accretion flow \citep{Tchekhovskoy2011, McKinney2012}. As it has been theoretically estimated and numerically confirmed, the spin-extraction/MAD scenario allows launching jets with the power up to $\simeq\! 3 \dot M c^2$ or so. The applicability of the spin-extraction/MAD model to radio loud quasars was questioned by \citet{vanVelzen2013}, who, by using energetics of radio lobes, found that the median value of the jet production efficiency, $\eta_{\rm j} \equiv P_{\rm j}/\dot M c^2$, in a radio-selected sample of quasars with double radio sources is $\sim$0.01. If $\eta_{\rm j}$ depended only on the BH spin, this would imply its average value much lower than those predicted by BH cosmological evolution \citep{Volonteri2013}. On the other hand, large efficiencies, with $\eta_{\rm j}\sim 0.1$--10, have been found using blazar models (\citealt{Ghisellini14}, hereafter G14). Similar efficiencies were then obtained by means of core-shift measurements in radio-core dominated quasars (\citealt{Zamaninasab14}, hereafter Z14; \citealt{Zdziarski15}, hereafter Z15). In this paper, we compare results obtained by all of these methods. In order to minimize effects of selection biases in comparison of different samples, we use cross-matched sub-samples. The paper is organized as follows. We define our observational samples in Section~\ref{samples}. Section~\ref{methods} presents the methods used to calculate the jet and accretion powers. Section~\ref{results} presents our results, which are then discussed in Section~\ref{discussion}.

1607.08895

1607

1607.04133_arXiv.txt

One of the most enigmatic and hitherto unexplained properties of Jupiter Trojans is their bimodal color distribution. This bimodality is indicative of two sub-populations within the Trojans, which have distinct size distributions. In this paper, we present a simple, plausible hypothesis for the origin and evolution of the two Trojan color sub-populations. In the framework of dynamical instability models of early Solar System evolution, which suggest a common primordial progenitor population for both Trojans and Kuiper belt objects, we use observational constraints to assert that the color bimodalities evident in both minor body populations developed within the primordial population prior to the onset of instability. We show that, beginning with an initial composition of rock and ices, location-dependent volatile loss through sublimation in this primordial population could have led to sharp changes in the surface composition with heliocentric distance. We propose that the depletion or retention of H$_{2}$S ice on the surface of these objects was the key factor in creating an initial color bimodality. Objects that retained H$_{2}$S on their surfaces developed characteristically redder colors upon irradiation than those that did not. After the bodies from the primordial population were scattered and emplaced into their current positions, they preserved this primordial color bimodality to the present day. We explore predictions of the volatile loss model --- in particular, the effect of collisions within the Trojan population on the size distributions of the two sub-populations --- and propose further experimental and observational tests of our hypothesis

In the past decade, Jupiter Trojans have been the subject of increasing scientific interest, and understanding their origin and surface properties promises to unlock many of the fundamental aspects of Solar System formation and evolution. These minor bodies share Jupiter's orbit around the Sun at 5.2~AU and reside in two clusters at the stable L4 and L5 Lagrangian points. One of the most important discoveries about the physical properties of Trojans is the existence of two sub-populations --- the less-red (LR) and red (R) Trojans --- whose members differ categorically with respect to several photometric and spectroscopic properties. Bimodality has been reported in visible and near-infrared colors \citep{szabo,roig,emery}, as well as in longer wavelength reflectance \citep{grav}. The two sub-populations are present in both the L4 and L5 swarms. Meanwhile, members of the only robustly attested collisional family within the Trojans --- the Eurybates family at L4 \citep{broz} --- all belong to the less-red sub-population \citep{fornasier}, an observation that has important implications for the interplay between self-collisions and the surface properties of these objects. While recent spectroscopic study in the near-infrared has detected water ice on the surfaces of some large Trojans, along with weaker absorption features possibly attributable to organics \citep{brown2015}, the overall surface compositions of Trojans remain poorly constrained. In particular, the color bimodality has no current explanation. Although some theories have been forwarded to produce the range of Trojan colors through ongoing processes such as surface gardening and/or various irradiation mechanisms \citep[e.g.,][]{melita}, none of these theories predict a bimodality in color, instead yielding a single broad color distribution that is inconsistent with observations. On the basis of the differing spectral properties of LR and R Trojans,\citet{emery} suggested that perhaps one of the two sub-populations formed in the Main Belt asteroid region, while the other formed in the outer Solar System before being captured by Jupiter. Recent theories of Solar System evolution, however, suggest that after an early dynamical instability between giant planets, the Jupiter Trojan region was filled exclusively with objects that formed at large heliocentric distances within an extended planetesimal disk situated beyond the primordial orbits of the ice giants \citep{morbidelli,roig2}. In the context of these dynamical instability models, both LR and R Trojans must have originated from the same primordial population. To date, no hypothesis has been proposed to explain the existence of two distinct sub-populations from a single source region. Support for the idea of a shared formation environment for the Trojans and Kuiper belt objects (KBOs) comes from the observation that the size distributions of Trojans and hot (i.e., dynamically excited) KBOs are consistent with each other \citep{fraser2014}. In addition, the total Trojan size distribution for objects smaller than $\sim$100~km is consistent with collisional equilibrium \citep[e.g.,][]{marzari}. Unexpectedly, however, the size distributions of the individual LR and R Trojan sub-populations are highly distinct; specifically, the power-law slopes of the magnitude distributions for objects smaller than $\sim$100~km in diameter are discrepant, revealing a monotonic increase in the relative number of LR Trojans with decreasing size \citep{wong,wong2}. Such distinct size distributions are difficult to reconcile with the hypothesis that the Trojans share a single source population. However, \citet{wong} demonstrated that, starting from an initial state where both sub-populations had identical magnitude distributions, a process wherein the collisional fragments of both R and LR Trojans become LR objects would naturally account for the relative depletion of R Trojans and the simultaneous enrichment of LR Trojans with decreasing size, as well as the collisional equilibrium of the overall population. While this explanation did not appeal to specific chemical or physical processes for the color conversion, the consistency between simulated and observed magnitude distributions suggests that R-to-LR conversion may provide a promising angle from which to address the question of the Trojans' color bimodality and the different LR and R size distributions. In this paper, we propose a simple, chemically plausible hypothesis, developed within the framework of current dynamical instability models, to explain the origin of the color bimodality within the Trojans and the differing size distributions of the two color sub-populations. Central to our hypothesis is location-dependent volatile loss, which established sharp changes in \textit{surface} composition and, thus, color across the primordial trans-Neptunian planetesimal disk prior to the onset of dynamical instability. In particular, we suggest that the loss or retention of H$_{2}$S due to sublimation resulted in two types of surface chemistry among objects in the primordial disk, subsequently leading to different characteristic colors upon irradiation. In Section~\ref{sec:obs}, we briefly summarize the present understanding of both Trojan and KBO colors, which serves to motivate and constrain our hypothesis. The location-dependent volatile loss model is detailed in Section~\ref{sec:volatiles}, and finally, the predictions of our hypothesis with respect to surface colors, collisions, and size distributions are discussed in Section~\ref{sec:colors}.

We have proposed a simple hypothesis for the two color sub-populations within the Trojans. The main feature of our proposal is location-dependent sublimation of volatile ices within a primordial planetesimal disk in the outer Solar System, which divided objects in this region into two groups --- those that retained H$_{2}$S on their surfaces, and those that did not. As a result of irradiation chemistry on these initially pristine icy surfaces, these two groups developed different involatile surface colors, thereby imprinting an initial color bimodality into the population. The retention of H$_{2}$S on the surface acted as a significant reddening agent upon irradiation, leading to characteristically redder colors than in the face of H$_{2}$S-depleted surfaces. Following a period of dynamical instability, the objects in this primordial disk were scattered throughout the middle and outer Solar System, carrying the primordial color bimodality with them as they populated the Trojan region and the current Kuiper belt. Our hypothesis thus offers a chemically and dynamically plausible explanation for the color bimodality of both Trojans and KBOs, as summarized in Figure~\ref{diagram}. In addition, our model predicts that all collisional fragments in the Trojan swarms are compositionally identical and become relatively less-red upon irradiation, regardless of whether the parent body is a LR or a R object. Therefore, our hypothesis accounts for the relative depletion of R objects with decreasing size and offers a self-consistent explanation for the discrepancy between the observed LR and R Trojan size distributions. Although it is supported by observational constraints and some experimental studies, the hypothesis advanced here is necessarily speculative. The most fruitful pathway for follow-up study would be systematic laboratory experiments that analyze the effects of irradiation on the color and reflectivity of surfaces analogous to the ones described in our hypothesis --- water ice-dominated mixtures of methanol, ammonia, and carbon dioxide ice, with or without the addition of H$_{2}$S ice. Crucially, these experiments must simulate the changing environments of Trojans and KBOs as they evolve within the primordial planetesimal disk and then are emplaced in their respective current locations. The main results that must be obtained in the experimental context for our hypothesis to be true are: (1) Irradiation of ice mixtures with methanol, ammonia, and carbon dioxide, but without H$_{2}$S, leads to reduction of albedo and a surface residue that has a color comparable to those seen on R KBOs and Centaurs. (2) Irradiation of ice mixtures with the addition of H$_{2}$S leads to reduction of albedo and a significantly redder irradiation mantle than in the absence of H$_{2}$S. (3) Intensified irradiation and higher temperatures alter the surface colors to those characteristic of Trojans. While some of these results are supported by laboratory data in the literature, only experimental verification using a uniform setup and conditions appropriate to the various minor body populations will allow us to confirm or refute the hypothesis for the color bimodality of Trojans presented here.

1607.04133

1607

1607.04305_arXiv.txt

This paper presents Gemini-$gri'$ high quality photometry for cluster candidates in the field of $NGC~1316$ (Fornax A) as part of a study that also includes GMOS spectroscopy. A preliminary discussion of the photometric data indicates the presence of four stellar cluster populations with distinctive features in terms of age, chemical abundance and spatial distribution. Two of them seem to be the usually old (metal poor and metal rich) populations typically found in elliptical galaxies. In turn, an intermediate-age (5 Gyr) globular cluster population is the dominant component of the sample (as reported by previous papers). We also find a younger cluster population with a tentative age of $\approx$ 1 Gyr.

\label{Intro} Once considered as `simple systems', Globular Clusters (GCs) are steadily leaving that characterization as more complex features of their stellar populations are discovered, e.g., \citet{Carretta2015}. This situation emphasizes the problem of not only understanding their formation as individuals but also in the context of galaxy formation \citep[e.g.][]{Brodie2014, Kruijssen2015, Kruijssen2016, Harris2015}.\\ The idea that GC systems are connected with large scale features of galaxies has its roots in \citet*{Eggen1962}. A recent example of this kind of analysis can be found in \citet{Forbes2016}.\\ If GCs are in fact tracers of the dominant stellar populations formed in different events during the life of a galaxy, they should reflect some common features with field stars (e.g., in terms of ages, chemical abundances and spatial distributions).\\ In this frame, $NGC~1316$, a giant elliptical galaxy and strong radio source (Fornax A), appears as a particularly attractive object. On one side, the galaxy displays a number of morphological features that seem the fingerprints of `merger' activity (shells, ripples, complex dust lanes) that have been studied in the optical range, for example, by \citet{Schweizer80, Schweizer81}. On the other, the galaxy exhibits a prominent GC system that has distinctive characteristics when compared with other bright ellipticals. The presence of `intermediate' age clusters in this galaxy, and their importance in the context of GCs formation, was already pointed out by \citet{Goudfrooij2001a} and subsequent studies \citep{Goudfrooij2001b, Goudfrooij2004, Goudfrooij2012}.\\ A key feature in this analysis is the identification of the different kind of cluster systems that co-exist in $NGC~1316$. For example, \citet{Goudfrooij2001b} show that the integrated brightness-colour domain occupied by the `blue' GCs in this galaxy is very similar to that of the low metallicity and old halo clusters in the Milky Way (MW). In turn, that work also pointed out that `intermediate' colour GCs are considerably brighter in the average, then suggesting younger ages.\\ Some of the features of the $NGC~1316$ GC system were studied by \citet{Gomez2001} on the basis of $BVI$ photometry. That work derived an overall ellipticity of 0.38 with a position angle of 63 degrees for the whole cluster system. Besides, their analysis of the projected areal density of the GCs as a function of galactocentric radius suggests that both the `blue' and `red' subpopulations share, to within the errors, very similar slopes.\\ A more recent attempt to disentangle the GCs populations using wide-field $(C-R)$ photometry has been presented by \citet{Richtler2012b}. These authors explore the possible presence of two or three cluster populations and conclude that the dominant component is as young as 2 Gyr.\\ In turn, \citet{Richtler2014} presented a thorough study of the kinematic behaviour of the GC system of $NGC~1316$ based on the radial velocities of 177 GCs. In another work, \citet{Richtler2012a} concentrated on the so called SH2 object, finding that it is in fact an unusual region of star formation.\\ A recent estimate of the distance modulus of $NGC~1316$ has been presented by \citet{Cantiello2013} who derive $(m-M)_{0}=31.59$ (20.8 Mpc) by means of the SBF method, that we adopt in this paper, and is somewhat smaller than that given in \citet{Goudfrooij2001a}: $(m-M)_{0}$=31.80.\\ In this work we present high quality Gemini $gri'$ photometry carried out on a CCD mosaic including eight different fields. This material is part of a study (in progress) that also includes GMOS spectroscopy for some 35 confirmed GCs. Currently, deep GC spectroscopic data is only available for three GCs \citep{Goudfrooij2001a}.\\ Low photometry errors are crucial for a characterization of the different cluster populations. In turn, the use of three different filters allows to account for field contamination and also for a comparison with simple stellar population (SSP) models in two colour diagrams.\\ The paper is organized as follows: The characteristics of the data handling and photometry, including an analysis of the errors and completeness, is presented in Section \ref{sec2}. At the same time, in this section we made the selection of unresolved sources and examined their colour-magnitude diagram. In Section \ref{sec3} we analyzed the distribution of the GCs on the sky as well as the colour-colour relations. An attempt to derive ages and chemical abundances using SSP models is presented in Section \ref{sec4}. The spatial distribution of each GC subpopulation and the behaviour of the areal density profiles is discussed in Section \ref{sec5}. A brief discussion of the radial velocities and of the GC integrated luminosity function are given in Section \ref{sec6} and \ref{sec7}, respectively. Finally, a summary of the main conclusions is presented in Section \ref{sec8}.\\

\label{sec8} The discussion of the $gri'$ photometry of cluster candidates in $NGC~1316$ presented in this paper, confirms earlier results that pointed out the complexity of the cluster system in this galaxy.\\ We identify four cluster subpopulations with distinct characteristics, namely:\\ a) Very blue clusters. \noindent These cluster candidates occupy a $(g-i)'_{0}$ colour range from 0.30 to 0.75. On one side, objects brighter than $g'_{0}=$23.5, show a very sparce distribution. Most of them show a colour distribution compatible with a 1 Gyr isochrone. Given the lack of unicity in the colour-abundance relation for that age, the situation remains largely undetermined and requires a spectroscopic analysis to clarify the nature of these objects.\\ GC number $119$ in \citet{Goudfrooij2001b}, falls in that colour range and is a confirmed member of the $NGC~1316$ system, on the basis of its radial velocity. This cluster has been identified by \citet{Richtler2012b} as a presumably very young cluster.\\ These bright cluster candidates are an intriguing subpopulation that, if confirmed as such, would indicate the existence of an intense burst of cluster formation widespread on the whole body of the galaxy.\\ In turn, fainter candidates within the same colour range, show a marked increase of the areal density in an annular region (60 to 120 arcsecs) around the center of the galaxy. The relatively large errors of the photometry for these faint clusters, prevent a significant analysis of their position in the two colour diagram. Their exponential luminosity function suggests that they might rather be massive young clusters than GCs.\\ b) Blue GCs. \noindent This subpopulation exhibits a colour peak at $(g-i)'_{0}$=0.82, similar to those observed for the blue GCs in elliptical galaxies. As noticed by \citet{Goudfrooij2001b} their photometric features are compatible with the old low metallicity GCs in the MW halo.\\ c) Intermediate GCs. \noindent The colours of these clusters correspond to a 5 Gyr old population in the frame of the Bressan et al. models. This age, and the mean chemical abundance we derive for these clusters (slighthly subsolar), are comparable with the spectroscopic age (3 $\pm$ 1 Gyr) and abundance (Solar) obtained by \citet{Goudfrooij2012} for three GCs which, according to their colours, belong to this population.\\ The spatial distribution of the intermediate GCs exibits a flattenig $q$ $\approx$ 0.7 and a position angle of $63$ degrees, i.e., some 8 degrees larger than that of the semimajor axis of the $NGC~1316$ halo. \citet{Gomez2001} find the same position angle for what they call `red' GCs. Their areal density profile can be fit by S\'ersic profile with $n=$ 1 (disc-like). \\ An analysis of the radial velocities of 28 intermediate GCs (with data from \citealt{Richtler2014}) shows marked similitude with the kinematics of the stellar halo of the galaxy. This result argues in favor of a connection between the intermediate GCs and field stars. The presence of an intermediate age stellar population was pointed out by \citet{Cantiello2013} through a comparison of their SBF colours and SSP model colours.\\ d) Red GCs. \noindent The colours of this population are compatible with those of old red GCs in elliptical galaxies. Their spatial distribution, spheroidal and slightly flattened, is coherent with a bulge-like high chemical abundance population (somewhat less than that of the intermediate GCs). They seem clearly distinct from the intermediate clusters. If the red GCs were coeval with the intermediate clusters, their chemical abundance should be three to four times larger than those we find on the basis of the model fits, not a very likely situation.\\ A summary of all these features then indicates the existence of a rather spherical low metallicity halo, and of a more chemically enriched and flattened bulge, coexisting with an `intermediate age' flattened spheroid (or even a thick disc?) which exhibits photometric and kinematic similarities with the galaxy halo. This scenario is compatible with \citet{McNeil2012} who suggest that $NGC~1316$ may represent the early stages of a system that would evolve to become a `Sombrero' like galaxy ($NGC~4594$) through a series of mergers. The so called `very blue' clusters, for which we find a tentative age of 1 Gyr, may be the tracers of the last of these events.\\

1607.04305

1607

1607.02379_arXiv.txt

{We study the propagation of a fast magnetoacoustic wave in a 3D magnetic field created from two magnetic dipoles. The magnetic topology contains an X-line.} {We aim to contribute to the overall understanding of MHD wave propagation within inhomogeneous media, specifically around X-lines.} {We investigate the linearised, 3D MHD equations under the assumptions of ideal and cold plasma. We utilise the WKB approximation and Charpit's method during our investigation.} {It is found that the behaviour of the fast magnetoacoustic wave is entirely dictated by the local, inhomogeneous, equilibrium Alfv\'en speed profile. All parts of the wave experience refraction during propagation, where the magnitude of the refraction effect depends on the location of an individual wave element within the inhomogeneous magnetic field. The X-line, along which the Alfv\'en speed is identically zero, acts as a focus for the refraction effect. There are two main types of wave behaviour: part of the wave is either trapped by the X-line or escapes the system, and there exists a critical starting region around the X-line that divides these two types of behaviour. For the set-up investigated, it is found that $15.5\%$ of the fast wave energy is trapped by the X-line.} {We conclude that linear, $\beta=0$ fast magnetoacoustic waves can accumulate along X-lines and thus these will be specific locations of fast wave energy deposition and thus preferential heating. The work here highlights the importance of understanding the magnetic topology of a system. We also demonstrate how the 3D WKB technique described in this paper can be applied to other magnetic configurations.}

It is now clear that magnetohydrodynamic (MHD) wave motions (e.g. Roberts \cite{Bernie}; Nakariakov \& Verwichte \cite{NV2005}; De Moortel \cite{DeMoortel2005}) are ubiquitous throughout the solar atmosphere (Tomczyk et al. \cite{Tomczyk}). Several different types of MHD wave motions have been observed by various solar instruments: longitudinal propagating disturbances have been seen in {{SOHO}} data (e.g. Berghmans \& Clette \cite{Berghmans1999}; Kliem {{et al.}} \cite{Kliem}; Wang {{et al.}} \cite{Wang2002}) and {{TRACE}} data (De Moortel {{et al.}} \cite{DeMoortel2000}) and these have been interpreted as slow magnetoacoustic waves. Transverse waves have been observed in the corona and chromosphere with {{TRACE}} (Aschwanden {{et al.}} \cite{Aschwandenetal1999}, \cite{Aschwandenetal2002}; Nakariakov {{et al.}} \cite{Nakariakov1999}; Wang \& Solanki \cite{Wang2004}), {{Hinode}} (Okamoto {{et al.}} \cite{Okamoto}; De Pontieu {{et al.}} \cite{Bart2007}; Ofman \& Wang \cite{OW2008}), SDO data (e.g. McIntosh et al. \cite{McIntosh2011}; Morton et al. \cite{Morton2012}, \cite{Morton2015}; Morton \& McLaughlin \cite{MortonMcLaughlin2013}, \cite{MortonMcLaughlin2014}; Thurgood et al. \cite{Thurgood2014}) and these have been interpreted as fast magnetoacoustic waves, specifically kink waves. These transverse motions have also been interpreted as Alfv\'enic waves, although this interpretation is subject to discussion, e.g. see Erd{\'e}lyi \& Fedun (\cite{RF2007}), Van Doorsselaere {{et al.}} (\cite{Tom2008}) and Goossens et al. (\cite{Goossens2009}). Non-thermal line broadening due to torsional Alfv\'en waves has been reported by {Erd{\'e}lyi} {{et al.}} (\cite{E1998}), Harrison {{et al.}} (\cite{Harrison2002}), O'Shea {{et al.}} (\cite{Oshea}) and Jess {{et al.}} (\cite{Jess2009}). It is also clear that the coronal magnetic field plays a fundamental role in the propagation and properties of MHD waves, and to begin to understand this inhomogeneous magnetised environment it is useful to look at the topology (structure) of the magnetic field itself. Potential-field extrapolations of the coronal magnetic field can be made from photospheric magnetograms (e.g. see R{\'e}gnier \cite{Stephane2013}) and such extrapolations show the existence of important features of the topology: {\emph{null points}} - specific points where the magnetic field is zero, {\emph{separatrices}} - topological features that separate regions of different magnetic flux connectivity, and {\emph{X-lines}} or {\emph{null lines}} - extended locations where the magnetic field, and thus the Alfv\'en speed, is zero. Investigations of the coronal magnetic field using such potential field calculations can be found in, e.g., Brown \& Priest (\cite{BrownPriest2001}), Beveridge et al. (\cite{Beveridge2002}), R{\'e}gnier et al. (\cite{Stephane2008}) and in a comprehensive review by Longcope (\cite{L2005}). These two areas of scientific study, namely ubiquitous MHD waves and magnetic topology, will naturally encounter each other in the solar atmosphere, e.g. MHD waves will propagate into the neighbourhood of coronal null points, X-lines and separatrices. Thus, the study of MHD waves within inhomogeneous magnetic media is itself a fundamental physical process. Previous works, detailed below, have focused on MHD wave behaviour in the neighbourhood of null points and separatrices (see review by McLaughlin et al. \cite{McLaughlinREVIEW}). However, less attention has been given to the transient behaviour of MHD waves in the vicinity of X-lines in the solar atmosphere. The motivation for this paper is to address this, i.e. this paper aims to investigate the behaviour of fast MHD waves around an X-line in order to contribute to the overall understanding of MHD wave propagation within inhomogeneous media. {{Note that an X-line is a degenerate structure and its existence requires a special symmetry of the magnetic field. Thus, given the inherent lack of symmetry in solar magnetic observations, their existence in the solar atmosphere is unlikely. However, }} X-lines are well studied in other areas, such as in the Earth's magnetosphere, e.g. Runov et al. (\cite{Runov2003}) and Phan et al. (\cite{Phan2006}). The propagation of fast magnetoacoustic waves in an inhomogeneous coronal plasma has been investigated by Nakariakov \& Roberts (\cite{Nakariakov1995}), who showed that the waves are refracted into regions of low Alfv\'en speed (see also Thurgood \& McLaughlin \cite{Thurgood}). In the case of X-lines, the Alfv\'en speed actually drops to zero. MHD waves in the neighbourhood of a single 2D X-point have been investigated by various authors. Bulanov \& Syrovatskii (\cite{Bulanov1980}) provided a detailed discussion of the propagation of fast and Alfv\'en waves using cylindrical symmetry. Craig \& Watson (\cite{CraigWatson1992}) mainly considered the radial propagation of the $m=0$ mode (where $m$ is the azimuthal wavenumber) using a mixture of analytical and numerical solutions. They showed that the propagation of the $m=0$ wave towards the null point generates an exponentially large increase in the current density. Craig \& McClymont (\cite{CraigMcClymont1991}, \cite{CraigMcClymont1993}), Hassam (\cite{Hassam1992}) and Ofman et al. (\cite{OMS1993}) investigated the normal mode solutions for both $m=0$ and $m\ne 0$ modes with resistivity included. They emphasise that the current builds as the inverse square of the radial distance from the X-point. All these investigations were carried out using cylindrical models in which the generated waves encircled the X-point and so the cylindrical symmetry meant that the disturbances can only propagate either towards or away from the X-point. The behaviour of MHD waves around two-dimensional X-points in a Cartesian geometry has been investigated by McLaughlin \& Hood (\cite{MH2004}, \cite{MH2005}, \cite{MH2006b}), McLaughlin et al. (\cite{MDHB2009}) and more recently by {Ku{\'z}ma} et al. ({\cite{Kuzma2015}). Of note is also McLaughlin \& Hood (\cite{MH2006a}) who investigated fast MHD wave propagation in the neighbourhood of two dipoles. These authors solved the linearised, $\beta=0$ MHD equations and found that the propagation of the linear fast wave is dictated by the Alfv\'en speed profile and that close to the X-point, the wave is attracted to the X-point by a refraction effect. It was also found that in this magnetic configuration a proportion of the wave can escape the refraction effect and that the split occurs near the regions of very high Alfv\'en speed. However, this study was limited to 2D. The current paper extends this work to 3D. MHD waves in the vicinity of a 3D null point (e.g. Parnell et al. \cite{Parnell1996}; Priest \& Forbes \cite{PF2000}) have also been investigated. Galsgaard et al. (\cite{Galsgaard2003}) performed numerical experiments on the effect of twisting the spine of a 3D null point, and described the resultant wave propagation towards the null. They found that when the fieldlines around the spine are perturbed in a rotationally symmetric manner, a twist wave (essentially an Alfv\'en wave) propagates towards the null along the fieldlines. Whilst this Alfv\'en wave spreads out as the null is approached, a fast-mode wave focuses on the null point and wraps around it. In addition, Pontin \& Galsgaard (\cite{PG2007}) and Pontin et al. (\cite{PBG2007}) performed numerical simulations in which the spine and fan of a 3D null point are subject to rotational and shear perturbations. They found that rotations of the fan plane lead to current sheets in the location of the spine and rotations about the spine lead to current sheets in the fan. The WKB approximation is an asymptotic approximation technique which can be used when a system contains a large parameter (see {{e.g.}} Bender \& Orszag \cite{Bender1978}). Hence, the WKB method can be used in a system where a wave propagates through a background medium which varies on some spatial scale which is much longer than the wavelength of the wave. There are several examples of authors utilising the WKB approximation to compare with numerical results, e.g. Khomenko \& Collados (\cite{Khomenko2006}) and Afanasyev \& Uralov (\cite{Afanasyev2011}, \cite{Afanasyev2012}). Galsgaard et al. (\cite{Galsgaard2003}) compared their numerical results with a WKB approximation and found that, for the $\beta=0$ fast wave, the wavefront wraps around the null point as it contracts towards it. They perform their WKB approximation in cylindrical polar coordinates and thus their resultant equations are two-dimensional, since a simple 3D null point is essentially 2D in cylindrical coordinates. In contrast, this paper will solve the WKB equations for three Cartesian components, and thus we can solve for more general disturbances and more general boundary conditions. McLaughlin et al. (\cite{MFH2008}) utilised the WKB approximation to investigate MHD wave behaviour in the neighbourhood of a fully 3D null point. The authors utilised the WKB approximation to determine the transient properties of the fast and Alfv\'en modes in a linear, $\beta=0$ plasma regime. From these works, it has been demonstrated that the WKB approximation can provide a vital link between analytical and numerical work, and often provides the critical insight into understanding the physical results. This paper demonstrates the methodology of how to apply the WKB approximation to a general 3D magnetic field configuration. We believe that with the vast amount of 3D modelling currently being undertaken, applying this WKB technique in 3D will be very useful and beneficial to the MHD modelling community. This paper describes an investigation into the behaviour of fast MHD waves around an X-line using the WKB approximation. The paper has the following outline: In $\S\ref{SEC:1}$, the basic equations, linearisation and assumptions are described, including details of our equilibrium magnetic field. $\S\ref{WKBAPPROXIMATION}$ details the WKB technique utilised in this paper as well as its application to the fast wave. The results are given in $\S\ref{sec:results}$ and the conclusions and discussion are presented in $\S\ref{conclusion}$. There are multiple appendices (\ref{appendixB}, \ref{appendixA}, \ref{appendixC}) which complement the work in the core text. \begin{figure*}[t] \begin{center} \includegraphics[width=5.35in]{F1_LOWER_RESOLUTION.eps} \caption{$(a)$ Equilibrium magnetic field in the $y=0$, $xz-$plane. Dipoles are located at $x=\pm 0.5$. X-point is located at $x=0$ and $z=\sqrt{2}a=\sqrt{0.5}=0.707$. Red lines indicate the separatrices in this plane. $(b)$ 3D visualisation of the equilibrium magnetic field denoting the red separatrices along $y=0$ from $(a)$ and perpendicular to this the blue X-line along $x=0$ from $(c)$. Equilibrium magnetic field is rotationally symmetric about the $y=0$ axis and thus black curves denote the separatrices in the $xy-$plane at $z=0$. $(c)$ Equilibrium magnetic {{field}} shown in the $x=0$, $yz-$plane. Magnetic field is only in the $x-$direction, hence no arrows. Blue line denotes the X-line of the form $y^2+z^2=2a^2$. $(d)$ Plot of $B_x(0,r)$ where $r^2=y^2+z^2$. $B_x(0,r)$ changes sign at $r = \sqrt{2}a= \sqrt{0.5}=0.707$, i.e. at location of the X-line. Maximum of $d B_x(0,r) / dr$ occurs at $r=1$, where $B_x(0,r=1)= (4/5)^{5/2} = 0.5724$.} \label{Figure1} \end{center} \end{figure*} \begin{figure*}[t] \begin{center} \includegraphics[width=5.35in]{F2_LOWER_RESOLUTION.eps} \caption{$(a)$ Shaded surface of $v_A(x,y,z_0) = | {\bf{B}}_0(x,y,z_0)|$ in the $xy-$plane at $z_0=0.1$, with local maxima at $(x,y,z)=(\pm a,0,0)$, i.e. the dipoles' location. $(b)$ Colour contour of $v_A(x,0,z) = | {\bf{B}}_0(x,0,z)|$ in the $y=0$, $xz-$plane. Contour is colour coded: $0 \le v_A \le 0.3$ (white); $0.3 \le v_A \le 0.5724$ (green); $0.5724 \le v_A \le 2$ (yellow); $2 \le v_A \le 30$ (orange); $v_A \ge 30$ (black). Red lines indicate the separatrices in this plane. $(c)$ Colour contour of $v_A(0,y,z) = | {\bf{B}}_0(0,y,z)|$ in the $x=0$, $yz-$plane. Blue line indicates the X-line in this plane. Contour is colour coded in the same way as $(b)$. $(d)$ Plot of $|B_x(0,r)|=v_A(0,y,z)$ where $r^2+y^2+z^2$ and axis displays $0.5 \le r \le 2$. Colour coding corresponds to that of $(b)$ and $(c)$, except now black represents $|B_x(0,r)| \le 0.3$.} \label{Figure2} \end{center} \end{figure*}

This paper describes an investigation into the behaviour of fast magnetoacoustic waves in the neighbourhood of two magnetic dipoles, under the assumptions of ideal and cold plasma. We have demonstrated how the WKB approximation can be used to help solve the linearised MHD equations and we have utilised Charpit's method and a Runge-Kutta numerical scheme during our investigation. For the fast magnetoacoustic wave, we find that the wave speed is entirely dictated by the local equilibrium Alfv\'en speed profile. For individual elements generated on a lower plane, where here a wavefront was generated on the $xy-$plane at $z_0=0.2$, we find that all parts of the wave experience a refraction effect, i.e. a deviation towards regions of lower Alfv\'en speed, and that the magnitude of the refraction is different depending upon where a fluid element is in the magnetic field configuration. We find that there are two main types of wave behaviour: \begin{itemize} \item{Individual fluid elements can be trapped by the X-line, {{spiralling}} into the X-line due to the refraction effect. These individual elements terminate at the X-line. The wave speed decreases as an element approaches the X-line and the speed is identically zero at the X-line. Hence, it cannot be crossed.} \item{Individual fluid elements can escape the system, where elements closer to the X-line have their ray paths modified to a greater extent that those {{farther}} away.} \end{itemize} Thus, {\emph{there is a critical starting point that divides these two types of behaviour}}. We find that in the $xz-$plane along $y=0$, this critical starting point is $x_0 = x_{\rm{critical}}= \pm 0.3$, and in the the $yz-$plane along $x=0$, this critical starting point is $y_0 = y_{\rm{critical}}= \pm 0.85$. For starting positions along the lines $y_0 = \pm x_0$, it was found that the critical starting point was $|x_0| = 0.252$. We can also estimate the amount of wave energy trapped by the X-line. For the system studied here, the fraction captured by the X-line will depend upon the critical starting point that divides the particle paths into those that spiral into the X-line and those that escape, as well as the overall length of the domain. In this paper we have set $-L \leq x \leq L$, $-L \leq y \leq L$ and $z=z_0=0.2L$, where $L$ is the length of our lower boundary and $L=1$ under our non-dimensionalisation. Thus along the line $x=0$, $y_{\rm{critical}}= \pm 0.85$ and so $85\%$ of the wave energy is trapped by the X-line, whereas along $y=0$, $x_{\rm{critical}}= \pm 0.3$ and so $30\%$ is trapped. Thus, more wave energy is trapped from the wave along the X-line than across it. Similarly, it was found that the critical starting point along the lines $y_0 = \pm x_0$ was $|x_0| = 0.252$, which corresponds to $17.8 \%$ of the wave energy being trapped along a diagonal line of length $\sqrt{2}L$, or $25.2 \%$ being trapped along a radial line of $L=1$. Further investigation shows that there is a critical $\emph{area}$ surrounding the X-line, within which all wave elements and thus wave energy is trapped. This critical area corresponds to $0.618 \: L^2$ across $-L \leq x \leq L$, $-L \leq y \leq L$ and $z=z_0=0.2L$, which corresponds to $15.5\%$ of the wave energy generated across an area $4\: L^2$ being trapped by the X-line. {{This critical area is fixed for the magnetic topology}} and so the percentage trapped decreases as one increases the area of the initial wave considered. We have also limited our investigation to understanding the fast wave, but we could have also investigated the second root of equation (\ref{F1}), i.e. the equations governing the Alfv\'en wave behaviour. To do so, we would assume $\omega^2 \neq { | {\bf{B}}_0 |}^2 {| {\bf{k}} |} ^{2}$ and investigate the resultant equations. We have included such a derivation in Appendix \ref{appendixB} although a full investigation is outside the scope of this current paper. The 3D WKB technique described in this paper can also be applied to other magnetic configurations and we hope that this paper has illustrated the potential of exploiting the technique. In addition, it is possible to extend the work by dropping the cold plasma assumption. This will lead to a third root of equation (\ref{F1}) which will correspond to the behaviour of the slow magnetoacoustic wave. {{When the cold plasma assumption is dropped the fast wave speed will no longer be zero along the X-line, and thus the wave may pass through it. McLaughlin \& Hood (\cite{MH2006b}) investigated the behaviour of magnetoacoustic waves in a finite-$\beta$ plasma in the neighbourhood of a two-dimensional X-point. It was found that the fast wave could now pass through the X-point due to the non-zero sound speed and that a fraction of the incident fast wave was converted to a slow wave as the wave crossed the $\beta=1$ layer. }} {{There are}} some caveats concerning the method presented here, i.e. if modellers wish to compare their work with a WKB approximation, it is essential to know the limitations of such a method. Firstly, in linear 3D MHD, we would expect a coupling between the fast wave and the other wave types due to the geometry. However, under the WKB approximation presented here, the wave sees the field as locally uniform and so there is no coupling between the wave types. To include the coupling, one needs to include the next terms in the approximation, {{i.e.}} the work presented here only deals with the first-order terms of the WKB approximation. Secondly, note that the work here is valid strictly for high-frequency waves, since we took $\phi$ and hence $\omega=\partial \phi / \partial t$ to be a large parameter in the system. The extension to low frequency waves is considered in Weinberg (\cite{Weinberg1962}). {{In this paper}} we have found that the X-line acts as a focus for the refraction effect and that this {{refraction effect is a key feature of fast wave propagation}}. Since the Alfv\'en speed drops to identically zero at the X-line, mathematically the wave never reaches there, but physically the length scales (i.e. the distance between, say, the leading edge and trailing edge of a wave pulse and/or wave train) will rapidly decrease, indicating that all gradients, including current density, will increase at this location. In other words, {\emph{the fast wave, and thus all the fast wave energy, accumulates along the X-line}}. If even a small amount of resistivity was included in our system, ohmic heating will extract the energy from this location. {{ Thus, we deduce that {{X-lines will be specific locations of fast wave energy deposition and preferential heating}}. This highlights the importance of understanding the magnetic topology of a system and it is at these areas where preferential heating will occur. This paper specifically concerns itself with preferential heating at the X-line. However, it is important to note that these are not the only topological locations at which heat deposition is expected.}} {{ Finally, we note that an X-line is a degenerate structure and its existence requires a special symmetry of the field, and any arbitrarily small perturbation to this symmetric configuration will lead to a magnetic topology without a true X-line. The resulting new topology may exhibit a non-zero component of ${\bf{B}}$ all along the original X-line, which may manifest as a quasi-separatrix layer, or as one or multiple null points. Should the symmetry be broken and the topology changed, we expect that (i) should a quasi-separatrix layer manifest, then we would still get the extreme stretching described in this paper, since the quasi-separatrix layer would be a location of rapidly changing magnetic field connectivity, and hence all gradients, including current density, may increase at these locations. (ii) Should null points appear, then we would expect to recover the results of McLaughlin et al. (\cite{MFH2008}), who studied wave propagation around 3D null points. }} \begin{acknowledgement} D.L. Spoors acknowledges an Undergraduate Research Bursary from the Royal Astronomical Society. The authors acknowledge IDL support provided by STFC. J.A. McLaughlin acknowledges generous support from the Leverhulme Trust and this work was funded by a Leverhulme Trust Research Project Grant: RPG-2015-075. J.A. McLaughlin wishes to thank Alan Hood for insightful discussions and helpful suggestions regarding this paper. \end{acknowledgement}

1607.02379

1607

1607.02186_arXiv.txt

In this work, we study the merger of two neutron stars with a gravitational mass of $1.4\usk M_\odot$ each, employing the Shen-Horowitz-Teige equation of state. This equation of state is a corner case, allowing the formation of a stable neutron star with the given total baryonic mass of $3.03 \usk M_\odot$. We investigate in unprecedented detail the structure of the remnant, in particular the mass distribution, the thermal structure, and the rotation profile. We also compute fluid trajectories both inside the remnant and those relevant for the formation of the disk. We find a peanut-shaped fluid flow inside the remnant following a strong $m=2$ perturbation. Moreover, the flow is locally compressive, causing the appearance of dynamic hot spots. Further, we introduce new diagnostic measures that are easy to implement in numeric simulations and that allow to quantify mass and compactness of merger remnants in a well-defined way. As in previous studies of supra- and hypermassive stars, we find a remnant with a slowly rotating core and an outer envelope rotating at nearly Keplerian velocity. We compute a Tolman-Oppenheimer-Volkoff star model which agrees well with that of the remnant in the core, while the latter possesses extensive outer layers rotating close to Kepler velocity. Finally, we extract the gravitational wave signal and discuss the detectability with modern observatories. This study has implications for the interpretation of gravitational wave detections from the post-merger phase, and is relevant for short gamma-ray burst models.

\label{sec:intro} Binary neutron star (BNS) mergers are one of the most interesting astrophysical scenarios since their description requires both general relativity (GR) in the strong field regime as well as nuclear physics for matter above nuclear density, in particular the equation of state (EOS) and neutrino physics. BNS mergers are also promising candidates for multimessenger astronomy. First, they are among the most promising sources of gravitational waves for the detection with ground-based interferometers such as advanced LIGO \cite{Harry:2010:84006} and Virgo \cite{Virgo2015}. Second, they are thought to be the energy source of short gamma ray bursts (SGRBs) \cite{Berger2014}, which are frequently observed among the most energetic explosions known in the universe. Furthermore, these events represent the most likely explanation for the abundance of heavy elements in the universe, complementary to supernova explosions \cite{Thielemann2011, Wanajo2014}. The hot and neutron-rich ejecta form the merger process can indeed undergo rapid neutron capture nuclear reactions (r-process reactions) which produce the heavy elements. Moreover, the radioactive decay of these elements heats up the material and leads to a so-called kilonova (or macronova) signal \cite{Li:1998:L59, Kulkarni:2005:macronova-term, Metzger2012}, emerging mostly in the optical band days/weeks after merger. Candidates for such a kilonova have already been found \cite{Tanvir2013, Berger2013}. Finally, in addition to SGRBs and kilonova signals, BNS mergers may be accompanied by other electromagnetic counterparts, including late-time radio signals (e.g., \cite{Metzger2012}) and, if the remnant is not a black hole (BH) but a long-lived metastable (or stable) NS~\cite{Giacomazzo2013ApJ...771L..26G}, luminous signals powered by the remnant spin-down and peaking in the x-ray band \cite{Siegel:2016a,Siegel:2016b,Fan:2006:372L19} or at lower energies \cite{Yu2013,MetzgerPiro2014}. The recent direct detection of two gravitational wave (GW) events caused by binary BH mergers \cite{LIGO:BBHGW:2016,LIGO_GW151226} has raised expectations for a detection of GWs from BNS mergers in the near future \cite{Abadie2010}. Such a detection might include a relatively strong signal from the post-merger phase. While the post-merger signal of binary BH mergers always consists of a short BH ringdown, for a BNS merger it can be much more complex. Moreover, this signal offers a unique opportunity to constrain the NS EOS~\cite{Bauswein:2012:11101, Takami:2014:91104, Takami2015}. The outcome of a BNS merger depends largely on the total mass of the system: while heavy systems promptly collapse to a BH during merger, lighter systems form a metastable or even stable NS. Merger remnants can be classified based on their mass. A hypermassive NS has a mass exceeding the maximum mass of a uniform rotating star, and a supramassive NS is below that mass but above the maximum mass for a non-rotating configuration. Those masses are also used to predict the lifetime of the merger remnant. A hypermassive neutron star (HMNS) collapses within tens of milliseconds, as soon as the degree of differential rotation becomes insufficient to stabilize it. Above a certain threshold mass, however, the merger results in a prompt collapse without a HMNS phase \cite{Bauswein:2013:131101}. A supramassive neutron star (SMNS) can be supported by uniform rotation for much longer spindown timescales, and eventually collapses when the rotational support becomes too small. In the early post-merger phase ($<100$ ms) the associated loss of angular momentum can be dominated by GW emission, while magnetic braking likely dominates on longer timescales. The exact values of the masses dividing the above cases depend sensitively on the EOS. Hence, the presence of a post-merger signal could rule out all EOS for which the masses determined from the inspiral signal would lead to a prompt collapse. Moreover, the detection of a very long post-merger signal (although more difficult due to decaying amplitude), could even rule out EOS predicting a HMNS as the outcome, instead of a SMNS. In addition, the post-merger GW spectrum is typically dominated by a single peak around a few kilohertz, which also depends on the EOS. Measuring the corresponding frequency might allow for an additional constraint on the EOS \cite{Bauswein:2012:11101,Takami:2014:91104, Takami2015}. In addition to the post-merger GW signal, there are additional reasons to consider a BNS merger leading to a long-lived remnant NS as a very important case. First, the recent discovery of single NSs with a mass of $\sim2~M_\odot$ \cite{Demorest:2010:1081,Antoniadis:2013:448} combined with the fact that the expected distribution of progenitor masses in BNS mergers is peaked around $1.3-1.4~M_\odot$ \cite{Belczynski2008} leads to the conclusion that forming a SMNS is a likely possibility, and in some cases it might be possible to form even a stable NS \cite{Giacomazzo2013ApJ...771L..26G}. A second reason is that recent observations of SGRBs by Swift \cite{Gehrels2004} revealed a large fraction of events accompanied by long-lasting (from minutes to hours) x-ray afterglows, which suggests the presence of a continuous injection of energy from a long-lived central engine (e.g.,~\cite{Rowlinson2013}). If SGRBs are associated with BNS mergers, this would suggest the presence of a long-lived (supramassive or stable) NS, instead of the commonly assumed BH surrounded by an accretion disk. The main difficulty of this so-called ``magnetar model" is the fact that in this case the strong baryon pollution surrounding the merger site could choke the formation of a relativistic jet and thus prevent the SGRB itself. To overcome such limitations, an alternative ``time-reversal" scenario \cite{Ciolfi:2015:36} has been proposed, in which a SMNS survives for some time powering the long-lasting x-ray emission and eventually collapses to a BH and launches the relativistic jet. Because of the optical depth of the surrounding environment, the spin-down-powered x-ray emission is delayed and can still be observed after the SGRB (thus appearing as an ``afterglow"). The main challenge to this scenario is the formation of a massive disk surrounding the BH, when the latter is formed not shortly after merger, but minutes or hours later. Studying the rotation profile, the disk structure and the amount of ejecta surrounding the merger site in the case of a long-lived NS is therefore necessary in order to shed light on the different SGRB scenarios. Models of differential rotation in the remnant NS often assume a rapidly rotating core, with the rotation profile described by the so-called $j$-const law \cite{Baumgarte:2000:29}. In contrast, a recent study \cite{Kastaun:2015:064027} demonstrated that HMNS or SMNS can indeed be produced with a slowly rotating core. In the case of HMNSs, the collapse can be significantly delayed because part of the mass forms an extensive outer bulge rotating close to Kepler velocity, thus centrifugally supported. Further studies \cite{Endrizzi:2016:164001} found similar behavior for more models, including an unequal mass system. The rotation profile is important for the oscillation frequencies of the core and hence the radiated GW signals. Also, the lifetime of a HMNS stabilized by a rapidly rotating core could differ strongly from a HMNS stabilized by a rapidly rotating outer bulge. This should be taken into account when attempting to deduce the EOS from the duration of a postmerger GW signal. Moreover, the differential rotation profile may also have an effect on magnetic field amplification and electromagnetic emission, as seen in previous studies employing the $j$-const law~\cite{Siegel:2014:6}. Therefore a careful investigation of the remnant properties also has important applications to electromagnetic counterparts of GWs from BNS mergers and, possibly, to SGRBs. Beside rapid differential rotation and larger mass, another difference to isolated NS is that shock heating during merger requires the consideration of thermal effects. A common notion is that HMNS are also stabilized by thermal pressure. For nuclear physics EOS however, the temperatures reached during merger do not increase the maximum mass of spherical or uniformly rotating NSs strongly, and, somewhat counter-intuitively, can even decrease it \cite{Kaplan:2014:19,Galeazzi:2013:64009}. However, the merger remnants are not heated homogeneously. It is possible that under-densities caused by ``hot spots'' in the rapidly rotating remnant contribute to the GW signal, with a frequency determined by the rotation rate at the location of the hot spot. Unless the hot spots occur near a maximum of the rotation rate, they would quickly turn into tightly wound spirals because of differential rotation, at which point the associated GW signal ceases. Because of this effect, both the exact nature of the fluid flow inside the remnant and the thermal structure, as well as their interplay, could become important for the interpretation of future GW detections. Note however that such contributions are likely smaller than the one from the non-axisymmetric oscillation of the remnant. Further, this effect competes with other secondary features caused by nonlinear combination frequencies between the radial and the main non-axisymmetric oscillation \cite{Stergioulas:2011:427} or side peaks caused by a strong frequency modulation of the main peak \cite{Kastaun:2015:064027}, again caused by radial oscillations. Even with a good signal to noise ratio for the post-merger phase, deriving information from secondary features of the post-merger spectrum will require a good theoretical understanding of all those effects. In this work, we will focus on a binary with typical mass but an EOS that allows the formation of a stable NS. This case is generically considered rather unlikely and for this reason it has not been investigated in detail. We note, however, that most of the results that will be presented in this study are also indicative for the SMNS case, which is instead a likely outcome of a BNS merger. We will investigate the structure of the fluid flow and thermal structure with regard to possible contributions to the post-merger GW signal. We will also investigate the rotation profile and mass distribution, and then we will take a closer look at the formation of the disk surrounding the remnant. Finally, we provide estimates for the mass ejection. Throughout this paper, we use units $G=c=1$ unless noted otherwise. We define baryonic mass $M_b = N_b m_b$ and rest mass density $\rho = n_b m_b$, where $N_b$ and $n_b$ are baryon number and baryon number density in the fluids restframe, respectively, and $m_b$ is a formal mass constant $m_b = 931.494 \usk\mega\electronvolt$.

\label{sec:summary} In this work, we investigated the merger of two NSs, each with a mass of $1.4 \usk M_\odot$, employing the Shen-Horowitz-Teige EOS. Because of the atypically large maximum mass allowed by this EOS, the result is a stable NS. We followed in detail the fluid flow during and after merger and found that for around $8$-$10\usk\milli\second$, the stellar cores remain roughly independent, rotating against each other while orbiting around each other. For the first ${\approx}3\usk\milli\second$, they are separated by a shock-heated layer with lower density. This layer is quickly ripped apart by the fluid flow, but the cores start merging only when the shear layer in between is dissipated into smaller vortices, with an increasing number of fluid trajectories enclosing both cores. Although our resolution is not sufficient to fully resolve the Kelvin-Helmholtz instability, we believe the qualitative picture is correct (even if the timescales might change). After the cores have merged, the fluid flow in the equatorial plane is peanut-shaped, following a strong $m=2$ perturbation. We studied the structure of the remnant near the end of the simulation, after it has settled down. The rotation profile has a maximum at a radius $15$--$20\usk\kilo\meter$, reaching a rotation rate around $960\usk\hertz$, while the core is slowly rotating with a rate of only $240\usk\hertz$ measured from infinity. Moreover, this rate is mostly due to frame dragging, and the core is almost nonrotating in the local inertial frame. The remnant has extended outer layers which approach Kepler velocity and smoothly join the disk orbiting the remnant. An important open question is about the prospect for the formation of a disk surrounding the BH even if the collapse happens much later than the merger, so that the original disk is already accreted onto the SMNS. This has only been investigated for simplified models with $j$-const rotation law or uniform rotation \cite{Baiotti:2007:187, Giacomazzo2011PhRvD..84b4022G, Margalit2015}. In this respect, a rotation profile like the one described above might or might not turn out to provide the necessary conditions. Of course, this would also require that the rapid rotation of the outer bulge of the remnant persists longer than the disk. This aspect will be further investigated in future work, and in particular in relation to the ``time-reversal" scenario for SGRBs \cite{Ciolfi:2015:36}, which envisages a late time collapse of the remnant into a BH plus accretion disk. To study the mass distribution of the remnant, we introduced a new measure that replaces density profiles, mass, and compactness in a way that can be used unambiguously for rapidly and differentially rotating merger remnants without a clearly defined surface. We found that the core has a structure very similar to the core of a specific TOV star. We set up both a cold TOV solution as well as one with the same average specific entropy as the remnant, and found that, although the hot star fitted the remnant better, thermal pressure has only a little effect on the average density profile of the core. Investigating the thermal evolution in more detail, we found that the average specific entropy of the remnant reaches $1 \usk k_B$ at $5\usk \milli\second$ after merger and then stays constant, while the disk is continuously heated up, reaching an average specific entropy $5\usk k_B$ at the end of the simulation. At this point, the average specific entropy at a given density raises from $0.3 \usk k_B$ at the center of the remnant to around $3 \usk k_B$ at the transition to the disk, and continues increasing toward lower densities. We note, however, that our simulation does not include neutrino radiation, which would likely change the temperature of the disk on the timescale of our evolution. One of our main interests was on the inhomogeneity of the remnant temperature. We found that quickly after merger the entropy of the remnant is concentrated in two hot spots which remain stable for around $15\usk\milli\second$. Those hot spots are locked in phase with the dominant $m=2$ perturbation, and consist partially of hot matter trapped in a small vortex in the ``waist'' of the peanut shaped remnant deformation. A significant part is however of dynamic nature, caused by adiabatic heating along a local compression. Because of the additional thermal pressure, it seems likely that the dynamic hot spot in turn has a backreaction on the remnant deformation and hence the fluid flow. The hot spots can thus alter the amplitude of the GW signal caused by the main deformation, but, being phase locked, do not contribute additional peaks to the power spectrum for the case at hand. We note, however, that in general secondary vortices such as the ones in our model might also undergo instabilities. Our study shows that the remnant structure is clearly affected by the vortices/hot spots; hence a rearrangement might slightly alter the remnant structure and hence lead to a sudden change of the GW frequency. The quantitative results we find are not general. Almost certainly results will change when considering unequal mass systems. Also the initial spin of the NSs has a strong influence, as we will show in a forthcoming publication. Nevertheless, we can conclude that the deformation of the remnant up to at least $20\usk\milli\second$ after merger should be treated as a complex nonlinear quasi-stationary state with relevant dynamic thermal effects. A state such as the one found in our work cannot be described well in terms of linear perturbations of axisymmetric rotating stars. Since the amplitude of the postmerger GW signal is decaying, this complicated period is highly relevant for the interpretation of future detections. Interpreting minor features of the GW spectrum such as side peaks in terms of oscillation mode analysis developed for isolated stars might be misleading. If a post-merger GW signal is detected, its duration can be taken as a lower limit for the lifetime of the remnant. In this regard it is relevant that a mounting number of results points to slowly rotating cores, even for HMNSs and SMNSs. The lifetime is then determined by the angular momentum balance of the outer layers and disk, not the core.

1607.02186

1607

1607.00018_arXiv.txt

Pulsars are some of the most accurate clocks found in nature, while black holes offer a unique arena for the study of quantum gravity. As such, pulsar--black hole (PSR--BH) binaries provide ideal astrophysical systems for detecting the effects of quantum gravity. With the success of aLIGO and the advent of instruments like the SKA and eLISA, the prospects for the discovery of such PSR--BH binaries are very promising. We argue that PSR--BH binaries can serve as ready-made testing grounds for proposed resolutions to the black hole information paradox. We propose using timing signals from a pulsar beam passing through the region near a black hole event horizon as a probe of quantum gravitational effects. In particular, we demonstrate that fluctuations of the geometry outside a black hole lead to an increase in the measured root mean square deviation of the arrival times of pulsar pulses traveling near the horizon. This allows for a clear observational test of the nonviolent nonlocality proposal for black hole information escape. For a series of pulses traversing the near-horizon region, this model predicts an rms in pulse arrival times of $\sim30\ \mu$s for a $3 M_\odot$ black hole, $\sim0.3\, $ms for a $30 M_\odot$ black hole, and $\sim40\, $s for Sgr A*. The current precision of pulse time-of-arrival measurements is sufficient to discern these rms fluctuations. This work is intended to motivate observational searches for PSR--BH systems as a means of testing models of quantum gravity.

Observable effects of quantum gravity are highly elusive. One reason is that the Planck scale, where such effects are expected to be manifest, is well beyond the reach of current experiments. In addition, regions of strong spacetime curvature, in which general relativity is expected to break down and be superseded by quantum gravity, tend to be hidden from view behind horizons. However, black holes may offer a unique low-energy window into quantum gravity. The process of black hole formation and evaporation poses a fundamental challenge to conventional low-energy physics. If unitarity is not violated in black hole evolution, information stored inside the black hole must somehow emerge and the local quantum field theoretic description of a semi-classical event horizon must be significantly modified. Quantum gravity will determine the way unitarity is ultimately preserved, and, if we are lucky, signals of this process may be detectable. It is currently an open question how quantum gravity resolves the black hole information paradox. Although most attempts to answer it have employed purely theoretical arguments, observational data may provide answers or at least constraints. While some alternatives, such as the firewall scenario \citep{Almheiri:2012rt}, predict phenomena that are extremely difficult to detect, other recent proposals feature large-scale nonlocal effects, e.g., \citet{Giddings:2004ud}, \citet{Papadodimas:2012aq}, \citet{Dvali:2012en}, \citet{Freidel:2015pka}, \citet{Hawking:2016msc}. In particular, the nonviolent nonlocality proposal of \citet{Giddings:2011ks, Giddings:2012gc, Giddings:2014nla, Giddings:2014ova, Giddings:2016tla} has potentially observable consequences. Thus, the practical question is, where is the best place to hunt for these elusive signals of quantum gravity? The Event Horizon Telescope (EHT) has generated substantial interest in the possibility of observing quantum gravitational effects in the near-horizon region of Sagittarius A* (Sgr A*), the supermassive black hole at the center of the Milky Way galaxy. Direct observations of the accretion disk might exhibit modification of structures, such as the black hole shadow and the photon ring, predicted by general relativity \citep{Giddings:2014ova}. % Another suggested method to investigate the near-horizon region, inspired by recent advanced Laser Interferometer Gravitational-wave Observatory (aLIGO) discoveries \citep{Abbott:2016blz, Abbott:2016nmj}, is to observe the gravitational waves emitted by the merger of two black holes. One could hope to detect deviations in the signal from the general relativistic prediction \citep{Giddings:2016tla}. The potential of gravitational wave astronomy has justifiably generated a great deal of excitement, but with so far only a handful of observed events so far, it is at best a promising but untested technique. Furthermore, both the substantial observational noise and the theoretical uncertainties inherent in the difficult numerical modeling of the inspiral process limit the ability to discern the effects of quantum gravity. Classically, the structure of spacetime is established by sending clock readings between observers on a spatial coordinate grid. To explore quantum mechanical effects on spacetime, we suggest using timing signals from the most precise natural clock---a pulsar---carried by light beams through the region near an event horizon. Therefore, we propose that a pulsar-black hole (PSR--BH) binary is an ideal astrophysical system for observing quantum gravitational effects. The fact that such a system is in some sense ``clean," that is, not muddled by other astrophysical features, allowed for precision measurements to be taken of a novel gravitational phenomenon, gravitational waves \citep{1982ApJ...253..908T,2010ApJ...722.1030W, 2016arXiv160602744W}. It is this critical property of such binary systems that allows them to be used as effective probes of quantum gravity. Pulsar--neutron star (PSR--NS) binaries are very clean systems whose orbital parameters can be measured with extreme accuracy, and for this reason, they have played a long and valuable role as precision astrophysical laboratories. Most famously, observations of the so-called ``binary pulsar'' (PSR 1913+16) were the first to confirm the existence of gravitational waves \citep{1982ApJ...253..908T}. In addition, general relativistic effects, such as the Shapiro time delay of pulsar radiation passing through the gravitational potential well of the companion NS, have been measured, yielding precision tests of GR \citep{2006Sci...314...97K}. PSR--BH binaries have been called the ``holy grail of astrophysics" \citep{FaucherGiguere:2010bq} both because of their unique potential and also because no actual examples have yet been found. However, the prospects for discovery are promising; large numbers of PSR--BH binaries are predicted to exist near the galactic center and should be readily detectable by both the Evolved Laser Interferometer Space Antenna (eLISA) \citep{AmaroSeoane:2012je} and the Square Kilometer Array (SKA) radio telescope \citep{FaucherGiguere:2010bq}. When they are eventually found, there are a variety of proposals to use PSR--BH binaries to investigate the gravitational properties of black holes \citep{2013ApJ...778..145N, Liu:2014uka, Rosa:2015hoa, 2016CQGra..33k3001J}, test extensions of general relativity \citep{Christian:2015smg, Yagi:2016jml}, and search for quantum gravitational effects associated with warped extra dimensions \citep{Simonetti:2010mk}. Additionally, \citet{2014MNRAS.445.3370P} and \citet{Pen:2014jqa} have suggested that lensing of pulsar emission passing near the horizon could provide increased sensitivity, allowing for the observation of quantum gravitational effects. For a pulsar orbiting a BH with an orbital plane that is seen edge-on, as the pulsar passes behind the BH, the radiation pulse travels through the near-horizon region. Because the pulsar signal can be characterized with exceptional precision, there is the possibility to detect even subtle quantum gravity effects. In particular, fluctuations of the near-horizon geometry alter the null geodesics along which the photons of the pulse travel, modifying the Shapiro time delay for each pulse. The effect can be observed as an increase in the root mean square (rms) variation in the arrival times of pulses at the telescope. By comparing pulse time-of-arrival (TOA) measurements when the pulsar is behind the BH to when it is in front, one can discern this increase in the rms due to these near-horizon quantum gravitational effects. We find that horizon-scale fluctuations of the near-horizon geometry of the magnitude estimated by \citet{Giddings:2014ova} will be detectable. For a series of pulses passing near the horizon, this version of the nonviolent nonlocality scenario predicts an rms in pulse arrival times of $\sim30$~$\mu$s for $3 M_\odot$ BH, $\sim0.3\, $ms for a $30 M_\odot$ BH, and $\sim40\, $s for Sgr A*. Standard radio-pulse TOA measurements have the ability to detect these effects. Observations of the type suggested here represent a definitive test of the model of nonviolent nonlocality presented in \citet{Giddings:2014nla}. This paper is organized as follows. Sec.~\ref{sec:info_paradox} explains why quantum gravitational effects outside a BH horizon might be observable in the context of resolving the BH information paradox. Sec.~\ref{sec:BH-NS_Binaries} provides more background on binary pulsars, particularly PSR--BH binaries, and Sec.~\ref{sec:PSR-Sgr_A*} considers the observational possibilities of pulsars in orbit around Sgr A*. Sec.~\ref{sec:time_delay} presents a theoretical calculation of the modification, due to metric fluctuations, of the TOA for pulses traveling near the BH, with a concrete example given in Sec.~\ref{sec:simpex}. Sec.~\ref{sec:Observables} explains how this effect can be observed. Finally, Sec.~\ref{sec:Discussion} concludes with open questions and directions for future research.

\label{sec:Discussion} In this paper we have argued that PSR--BH binaries are ideally suited to probe metric fluctuations near the event horizons of black holes. More specifically, we have demonstrated that a particular version of the nonviolent nonlocality proposal to resolve the BH information loss paradox predicts that these fluctuations are sufficiently large to generate an observable increase of the rms in pulsar TOA measurements, given by equation (\ref{eq:sigmadelta}). The required rate at which information must be released from the black hole implies $\kappa$ is of the order of one. It is thus important to note that $\kappa$ is not, in the context of this proposal, a tunable parameter. Our results demonstrate that observations of a PSR--BH binary with the right properties can be used to probe values of $\kappa\simeq1$, thus making such observations a definitive test of this scenario. Moreover, this method can be used to test other models of quantum gravity that predict anomalous behavior of the metric near the horizon. This current work could be extended in several ways. The relatively simple modification of the BH gravitational potential considered in Sec.~\ref{sec:simpex} could be made more detailed and sophisticated. In addition, numerical simulations could shed light on the precise effect such metric fluctuations would have on pulses traversing the near-horizon region. For a typical PSR--BH binary in which the pulsar and BH are well separated, the modification of the Shapiro time delay is an observable consequence of the metric fluctuations predicted by the nonviolent nonlocality proposal. However, if the pulsar itself travels through the near-horizon region, for example, in the final stages of a PSR--BH merger, other observable modifications to TOA measurements could arise. It is also worth considering ways in which PSR--BH binaries could be used to observe other quantum gravitational effects, such as enhanced Hawking flux or fluctuations in other metric components. Observational astronomy and theoretical work in quantum gravity have not traditionally had a great deal of overlap, except perhaps in a cosmological setting. However, Earth-based tests of quantum gravitational models are hard to pursue, while the universe provides ready-made astrophysical laboratories that can explore these extreme situations. In addition to providing high-precision tests of classical general relativity and other relativistic gravity theories, pulsar--black hole binaries provide testing grounds for aspects of quantum gravity. We hope this paper will lead to further exploration of these possibilities and encourage the search for such laboratories in the sky.

1607.00018

1607

1607.08623_arXiv.txt

Near-infrared color-excess and extinction ratios are essential for establishing the cosmic distance scale and probing the Galaxy, particularly when analyzing targets attenuated by significant dust. A robust determination of those ratios followed from leveraging new infrared observations from the VVV survey, wherein numerous bulge RR Lyrae and Type II Cepheids were discovered, in addition to $BVJHK_{s}(3.4\rightarrow22)\mu m$ data for classical Cepheids and O-stars occupying the broader Galaxy. The apparent optical color-excess ratios vary significantly with Galactic longitude ($\ell$), whereas the near-infrared results are comparatively constant with $\ell$ and Galactocentric distance ($\langle E(J-\overline{3.5\mu m})/E(J-K_s) \rangle =1.28\pm0.03$). The results derived imply that classical Cepheids and O-stars display separate optical trends ($R_{V,BV}$) with $\ell$, which appear to disfavor theories advocating a strict and marked decrease in dust size with increasing Galactocentric distance. The classical Cepheid, Type II Cepheid, and RR Lyrae variables are characterized by $\langle A_{J}/E(J-K_s) \rangle = \langle R_{J,JK_s} \rangle =1.49\pm0.05$ ($\langle A_{K_s}/A_J \rangle =0.33\pm0.02$), whereas the O-stars are expectedly impacted by emission beyond $3.6 \mu m$. The mean optical ratios characterizing classical Cepheids and O-stars are approximately $\langle R_{V,BV} \rangle \sim3.1$ and $\langle R_{V,BV} \rangle \sim3.3$, respectively.

Applying corrections for dust extinction is a ubiquitous task executed in disciplines throughout astronomy, yet key concerns persist regarding the topic. Cited color-excess and total-to-selective extinction ratios are contested \citep{tu12,na16}, and a debate continues regarding the compositional nature of dust and the source(s) behind diffuse interstellar absorption lines. The matter is exacerbated by the dependence of certain extinction ratios ($A_{V}/E(B-V)=R_{V,BV}$) on Galactic longitude $\ell$ \citep{sc16}, as exemplified by observations of the young cluster Westerlund 2 \citep[$\ell \sim 280 \degr$, $R_{V,BV}\sim 4$ versus $\langle R_{V,BV} \rangle \sim3.1$\footnote{\citet{ti14} obtained $\langle R_{V,BV} \rangle=2.40 \pm 1.05$ from high latitude SDSS BHB stars, whereas \citet{tu76} determined $\langle R_{V,BV} \rangle \sim3.1$ from open clusters.},][]{ca13}. Yet the challenge inherent in determining those ratios often requires the adoption of results tied to separate sight-lines and stellar populations, with a potential penalty being the propagation of systematic uncertainties. Caution is likewise warranted when aiming to subvert such difficulties by assuming a linear relationship between reddening and distance, since numerous sight-lines are characterized by non-linear trends \citep{nk80}. More broadly, \citet{pk12} and \citet{na16} argued that extinction laws adopted in surveys aiming to constrain cosmological models should be revisited \citep[e.g.,][]{ri11}, and there exist optimal passband combinations displaying less systematic and random scatter \citep[see also][]{ng12}. As a result of the aforementioned uncertainties, infrared observations are of particular importance when establishing the cosmic distance scale, as the wavelength regime exhibits a reduced sensitivity to dust obscuration relative to optical data (i.e., $A_{\lambda} \sim \lambda^{-\beta}$). Therefore, potential and often unconstrained variations in the extinction law are less onerous on the uncertainty budget (e.g., $\Delta \mu_0$). Spitzer observations of Cepheids throughout the Local Group exemplify that advantage \citep{sc11,ma13}, in concert with infrared monitoring of star clusters \citep{ch12,mb13}. For example, uncertainties associated with the reddening of the Large Magellanic Cloud, a pertinent anchor of the cosmic distance scale, are comparatively marginal in the mid-infrared where $\langle A_{3.6 \mu m}/E(B-V) \rangle \sim 0.18$ \citep{ma13}. An added advantage of infrared observations is the mitigated impact of compositional differences between calibrating and target standard candles when establishing distances, as line blanketing may affect optical $BV$ observations \citep{cc85,ma09}. In this study, reddening and total-to-selective extinction ratios (e.g., $A_{J}/E(J-K_s)$) are inferred from a diverse stellar demographic. That was accomplished by examining new $JHK_s$ observations from the $VVV$ survey \citep{mi10}, wherein RR Lyrae and Type II Cepheid variables were discovered toward the Galactic bulge and an adjacent region of the Galactic disk, in tandem with multiband infrared (e.g., Spitzer 3.6 $\mu m$) and optical data for O-stars and classical Cepheids throughout the Galaxy. The analysis aims to provide key insight regarding extinction, and assess whether the data corroborate findings implying a potential link between the Galactocentric distance and dust size, as indicated by red clump stars \citep[][discussion and references therein]{za09,go13}. \begin{figure}[!t] \begin{center} \includegraphics[width=8cm]{fig1.eps} \caption{Apparent color ratios for Type II Cepheids (top panel) and RR Lyrae (lower panel) variables identified in the $VVV$ and GLIMPSE surveys ($S_1\simeq 3.6 \mu m$). The stars are key for establishing a sizable Galactocentric baseline to evaluate reputed extinction ratio variations. For clarity purposes, only uncertainties along the ordinate are displayed.} \label{fig-cr} \end{center} \end{figure}

\label{s-conclusion} New near-infrared observations from the $VVV$ survey were employed to help establish color-excess and total extinction ratios across the Galaxy. To that end, Type II Cepheids and RR Lyrae variables were identified throughout the bulge and an adjacent region of the disk. The $VVV$ observations were paired with mid-infrared observations, and used to determine the desired ratios via apparent stellar colors (\S \ref{scolors}). The ensuing results were compared to color-excess ratios inferred from O-stars and classical Cepheids occupying the broader Milky Way. In sum, a strict paradigm linking smaller dust grains with increasing Galactocentric distance, as inferred from red clump stars, is not supported by the results. The optical analysis indicates that O-stars exhibit a maximum ratio ($R_{V,BV}$) along the $\ell \sim 290 \degr$ sight-line, and minima are located toward $\ell \sim 133 \degr$ and $\ell \sim 78 \degr$ (Fig.~\ref{fig-cel}). Conversely, the classical Cepheid optical data display extrema along the anti-centre and $\ell \sim 30 \degr$ sight-lines. Yet the infrared colors imply a ratio ($\langle E(J-\overline{3.5 \mu m})/E(J-K_s) \rangle =1.28\pm0.03 \sigma$) that is relatively constant in comparison to optical determinations (Figs.~\ref{fig-cel} \& \ref{fig-ceg}). The O-star color-excess ratios are particularly affected by emission beyond $3.6 \mu m$, however, the classical Cepheid, Type II Cepheid, and RR Lyrae variables may be characterized by $\langle R_{J,JK_s} \rangle =1.49\pm0.05$ (Fig.~\ref{fig-cer}), which implies $\langle A_{K_s}/A_J \rangle = 0.33\pm0.02$. The common distance approach yielded inconsistent (lower) results relative to those inferred from stellar colors for the bulge population. The former procedure is rather sensitive to the sample's distribution, and may be partially skewed by heavy extinction and blending (Fig.~\ref{fig-rr}). The mean optical total-to-selective extinction ratios are sensitive to $\ell$, but are approximately $\langle A_{V}/E(B-V) \rangle=\langle R_{V,BV} \rangle=3.1$ (classical Cepheids) and $\langle R_{V,BV} \rangle=3.3$ (O-stars). The ratios are consistent with the \citet{be96} and \citet{pk12} findings, and larger than the mean determined by \citet[][$R_{V,BV}=2.40 \pm 1.05$, from high $b$ SDSS BHB stars]{ti14}. Yet ultimately the $J,JK_s$ passband combination, and certain others \citep{ng12,na16}, present desirable advantages when establishing cosmic distances to Cepheids, which are used to constrain the Hubble constant, the Universe's age, and cosmological models \citep{fr96,ri11,ng13}. \subsection*{Acknowledgements} \scriptsize{D.M. (Majaess) is grateful to the following individuals and consortia whose efforts, advice, or encouragement enabled the research: 2MASS, WISE, GLIMPSE (Spitzer), L. Berdnikov, OGLE (A. Udalski, I. Soszynski), J. Ma{\'{\i}}z-Apell{\'a}niz, N. Walborn, C. Ngeow, L. Macri, D. Balam, B. Skiff, G. Carraro, CDS (F. Ochsenbein, T. Boch, P. Fernique), arXiv, and NASA ADS. D.M. (Minniti) is supported by FONDECYT Regular No. 1130196, the BASAL CATA Center for Astrophysics and Associated Technologies PFB-06, and the Ministry for the Economy, Development, and Tourism’s Programa Iniciativa Científica Milenio IC120009, awarded to the Millennium Institute of Astrophysics (MAS). W.G. gratefully acknowledges financial support for this work from the BASAL Centro de Astrofisica y Tecnologias Afines (CATA) PFB-06, and from the Millennium Institute of Astrophysics (MAS) of the Iniciativa Milenio del Ministerio de Economia, Fomento y Turismo de Chile, grant IC120009.}

1607.08623

1607

1607.08765_arXiv.txt

s{ Particles weakly interacting with ordinary matter, with an associated mass of the order of an atomic nucleus (WIMPs), are plausible candidates for Dark Matter. The direct detection of an elastic collision of a target nuclei induced by one of these WIMPs has to be discriminated from the signal produced by the neutrons, which leaves the same signal in a detector. The MIMAC (MIcro-tpc MAtrix of Chambers) collaboration has developed an original prototype detector which combines a large pixelated Micromegas coupled with a fast, self-triggering, electronics. Aspects of the two-chamber module in operation in the Modane Underground Laboratory are presented: calibration, characterization of the $^{222}$Rn progeny. A new test bench combining a MIMAC chamber with the COMIMAC portable quenching line has been set up to characterize the 3D tracks of low energy ions in the MIMAC gas mixture: the preliminary results thereof are presented. Future steps are briefly discussed. } \vspace{-6mm}

1607.08765

1607

1607.08848_arXiv.txt

We generalize the Rigid-Field Hydrodynamic equations to accommodate arbitrary magnetic field topologies, resulting in a new Arbitrary Rigid-Field Hydrodynamic (ARFHD) formalism. We undertake a critical point calculation of the steady-state ARFHD equations with a CAK-type radiative acceleration and determine the effects of a dipole magnetic field on the usual CAK mass-loss rate and velocity structure. Enforcing the proper optically-thin limit for the radiative line-acceleration is found to decrease both the mass-loss and wind acceleration, while rotation boosts both properties. We define optically-thin-correction and rotation parameters to quantify these effects on the global mass-loss rate and develop scaling laws for the surface mass-flux as a function of surface colatitude. These scaling laws are found to agree with previous laws derived from magnetohydrodynamic simulations of magnetospheres. The dipole magnetosphere velocity structure is found to differ from a global beta-velocity law, which contradicts a central assumption of the previously-developed XADM model of X-ray emission from magnetospheres.

\label{sec:Intro} In the last decade, spectropolarimetric surveys of OB stars have revealed that about 5-10\% of these massive stars have large-scale, organized magnetic fields (MiMeS: \citealp{wade14}; BOB: \citealp{morel15}). Such detectable magnetic fields ($B\gtrsim 100\unit{G}$) have a significant effect on the stellar wind, both channelling and trapping plasma within a stellar magnetosphere. This accumulated plasma produces extrastellar emission in optical (e.g. \citealp{howarth07}, \citealp{bohlender11}, \citealp{grunhut12} and references therein), infrared \citep{eikenberry14}, radio \citep{linsky92, chandra15}, and X-ray \citep{naze14, naze15}. Furthermore, this emission exhibits a rotational modulation as the plasma is forced by the magnetic field to co-rotate with the star. Similar advances in magnetosphere theory have also followed, starting with the pioneering magnetohydrodynamics (MHD) simulations of \citet{uddoula02}. They developed a ``wind magnetic confinement parameter'' to characterize the interplay between the stellar magnetic field and flow: \begin{align} \label{eq:eta_star}\eta_*\equiv \frac{B_\mathrm{eq}^2 R_*^2}{\dot M_{B=0} v_{\infty,B=0}}, \end{align} with $\dot M_{B=0}$ and $v_{\infty,B=0}$ being the stellar mass-loss rate and terminal velocity if the star had no magnetic field. The confinement parameter $\eta_*$ has become the canonical value adopted in scaling relations to explain the size \citep{uddoula02}, the mass-loss \citep{uddoula08}, the spin-down \citep{uddoula09}, and, with the critical rotation fraction $\omega$, the classification \citep{petit13} of magnetospheres. However, $\eta_*$ itself depends on non-magnetic values, ignoring any effects of the magnetic field. How does the magnetic field change the mass-loss rate and velocity? Can we use these new values to make a better confinement parameter? Traditionally, $\dot M$ and $v_\infty$ have been determined by analyzing the equation of motion for a line-driven wind (\citealp{cak75}; hereafter CAK) and solving for the so-called ``critical point''. Over the years, various modifications to the base CAK model (finite-disk effect: \citealp{friend86}, \citealp{pauldrach86}; depth-dependent force multiplier parameters: \citealp{kudritzki02}) have led to more realistic predictions of the mass-loss and terminal velocities. Other methods have been developed to improve on these estimates, such as a Monte Carlo method \citep{vink00, noebauer15} and a scattering source function technique \citep{sundqvist15}. For now, we use the CAK line-driving force in order to take the first steps towards understanding the effect of a dipole field on a stellar wind. In this paper, we present and study the Arbitrary Rigid-Field Hydrodynamics (ARFHD) equations, an extension of Rigid-Field Hydrodynamics (RFHD) \citep{townsend07} to account for non-dipole magnetic geometries (though we will consider only dipolar topologies in this analysis). RFHD was originally developed as an extension of the Rigidly Rotating Magnetosphere (RRM) model \citep{townsend05} for centrifugal magnetospheres, whose large magnetic fields make MHD simulations very impractical. In this ansatz, the magnetic fields are assumed to be completely rigid ($\eta_*\to\infty$), channeling the stellar wind along quasi-one-dimensional flux tubes. This allows each field line to be studied and simulated independently from one another, though this does miss important multi-dimensional effects present in the MHD simulations. In essence, the MHD studies approach the subject of massive-star magnetospheres from the regime of low magnetic confinement; ARFHD approaches this subject from the opposite regime of strong magnetic confinement. By blending both studies, we can set limits on the behavior of magnetospheres. In Section \ref{sec:hydro}, we present the reformulated ARFHD equations and define all the terms, including external sources of acceleration and cooling. Following this, we develop the critical point equations for an arbitrary magnetic configuration in Section \ref{sec:topo} and an algorithm for determing the critical point location in Section \ref{sec:crit}. Section \ref{sec:dip_topo} details the implementation and application of an aligned magnetic dipole radiation-driven wind model which includes the effect of stellar rotation. We present analytic scalings of the surface mass-flux in Section \ref{sec:mass_flux} and model results for the critical point location (Section \ref{sec:rcrit}), velocity structure (Section \ref{sec:term_veloc}), and, finally, the global mass-loss rate (Section \ref{sec:mass_loss}).

In this paper, we presented a critical point analysis of the Arbitrary Rigid-Field Hydrodynamic Equations, which represent a CAK-type wind within an arbitrary, infinitely-strong magnetic field. This differs from the usual CAK wind model by including the proper optically-thin maximum line-force, a rigid-body centrifugal acceleration, and a dipole areal divergence. After finding the general critical point values for the mass-flux, velocity and velocity derivative, we confirmed that they reduced to the proper values for a traditional CAK wind, i.e. a non-rotating, non-optically-thin corrected, radial flow with spherical divergence. These benchmarked general critical point equations were then applied to an aligned magnetic dipole field in order to calculate critical point locations and surface mass-fluxes. By integrating from these critical point locations, the velocity structure within the magnetosphere was quantified and studied. Finally, we obtained global mass-loss rates and found that the dipole field effectively reduces the overall mass-loss to 20-65\% of the non-magnetic, non-rotating CAK value. The key results are summarized as follows: \begin{enumerate} \item We are able to approximately confirm the \citet{owocki04} scaling for the influence of a magnetic dipole on the surface mass-flux, $\dot m_r\approx \mu_B^2\mcak{}$. While this scaling does not need much improvement, we provide a more accurate scaling equation (\autoref{eq:mflux_zero_unsat}) and detail which approximations are required to reproduce the OD04 scaling. \item The effect of a optically-thin corrected line-force can be encapsulated in a OTC parameter, which we call \sigsatz{} (\autoref{eq:sat_param}). Including this does not have much of an effect for O-type stars, since their increased wind density means that there will be less difference in the corrected and uncorrected line-forces. B-type stars, on the other hand, have their surface mass-flux reduced by approximately 25-30\% when the optically-thin correction is taken into account. \item The effect of rotation can be similarly represented with a rotation-effect parameter, which we call $\aleph$ (\autoref{eq:rot_param}). The amount of rotational boosting of the mass-flux is found to depend on both the rotational colatitude and the magnetic obliquity angle. \item The effects of rotation and the optically-thin correction can not be decoupled, however. We find a different OTC parameter in the case of rotation, \sigsatrot{} (\autoref{eq:sat_rot_param}). Rotation is found to reduce the correction by driving a higher surface mass-flux. \item The velocity structure within a magnetosphere cannot be described by a global beta-velocity law. However, at least for zero rotation, we can well-fit each line with individual beta-velocity laws. The best-fit $v_\infty$ and $\beta$ do vary from line to line, however. With rotation, the beta-velocity law assumption breaks down. \item The global mass-loss rate for a optically-thin corrected line-force can be accurately estimated by multiplying the optically-thick mass-loss by the OTC parameter, $\Sigma$. We find ``effective'' magnetospheric mass-loss rates, in which the plasma does not fall back to the star, to be approximately 20-65\% of the non-magnetic, non-rotating CAK mass-loss rate. \end{enumerate} Overall, we have quantified the effect of a magnetic dipole on a massive star wind with an eye towards better understanding of massive star magnetospheres. Next steps include adding the finite-disk correction parameter and quantifying its effect on the magnetospheric mass-loss and velocity (Paper II). Paper III will add colliding wind shocks and the subsequent ``cooling'' region to each line in order to better quantify the level of X-ray emission coming from each line. This will provide accurate initial conditions for hydrodynamical simulations of centrifugal magnetospheres. \begin{table*} \centering \caption{\label{tab:O_massloss}Same as \autoref{tab:B_massloss}, except for an O-type star with $\eta_* = 100$. All mass-loss rates are given in $10^{-6}~M_\odot$/yr. Numbers in parentheses next to a mass-loss rate represent the ratio of that particular rate to the ``General'' mass-loss with the same rotation.} \begin{tabular}{lccccccc} \hline\hline &No $B$ & $\omega = 0.0$ & 0.2 & 0.35 & 0.5 & 0.65 & 0.8 \\ \hline Optically-Thick & 6.61 & 3.77 & 3.79 & 3.83& 3.91 & 4.01 & 4.11\\ General & 6.44 & 3.67 & 3.70 & 3.74& 3.81 & 3.91 & 4.03\\ True & ... & 1.33(0.36) & 1.35(0.36) & 1.35(0.37) & 1.37(0.36) & 1.41(0.36)& 1.44(0.36) \\ Disk & ... & ... & ... & 0.52 (0.14) & 1.12 (0.29) & 1.78(0.46) & 2.56(0.64)\\ Effective & ... & 1.33 (0.36) & 1.35 (0.36) & 1.87(0.5) & 2.49(0.65) & 3.19(0.82) & 3.99 (0.99)\\ \hline \end{tabular} \end{table*}

1607.08848

1607

1607.06064_arXiv.txt

Advanced ACTPol is an instrument upgrade for the six-meter Atacama Cosmology Telescope (ACT) designed to measure the cosmic microwave background (CMB) temperature and polarization with arcminute-scale angular resolution. To achieve its science goals, Advanced ACTPol utilizes a larger readout multiplexing factor than any previous CMB experiment to measure detector arrays with approximately two thousand transition-edge sensor (TES) bolometers in each 150 mm detector wafer. We present the implementation and testing of the Advanced ACTPol time-division multiplexing readout architecture with a 64-row multiplexing factor. This includes testing of individual multichroic detector pixels and superconducting quantum interference device (SQUID) multiplexing chips as well as testing and optimizing of the integrated readout electronics. In particular, we describe the new automated multiplexing SQUID tuning procedure developed to select and optimize the thousands of SQUID parameters required to readout each Advanced ACTPol array. The multichroic detector pixels in each array use separate channels for each polarization and each of the two frequencies, such that four TESes must be read out per pixel. Challenges addressed include doubling the number of detectors per multiplexed readout channel compared to ACTPol and optimizing the Nyquist inductance to minimize detector and SQUID noise aliasing.

\label{sec:intro} Large arrays of low-temperature detectors are finding increasingly wider use in the detection of radiation, from photons in the sub-mm and gamma-rays, to neutrinos, and even dark matter. In Cosmic Microwave Background (CMB) and sub-mm astronomy, large arrays of detectors permit substantial improvements in sensitivity and mapping speed if the individual detectors are limited by the photon noise background. The most mature superconducting detector technology, the transition-edge sensor (TES)~\cite{Irwin95}, has demonstrated background limited performance across a range of bands and platforms. The present generation of ground-based CMB experiments (so-called Stage-III) plan to field roughly $10^{4}$ polarization sensitive detectors, and the CMB community is actively planning a future Stage-IV effort to field $10^{5}$-$10^{6}$ detectors~\cite{Wu14}, many of which may populate the focal planes of a small number of high-throughput telescopes~\cite{Niemack16}. Present Stage-III ground-based efforts include Advanced ACTPol (\AdvACT) on the six-meter Atacama Cosmology Telescope (ACT)~\cite{Henderson16}, BICEP3/Keck Array~\cite{Ogburn10,Staniszewski12}, CLASS~\cite{Essinger-Hileman14}, the Simons Array~\cite{Arnold14}, and SPT-3G on the South Pole Telescope~\cite{Benson14} (balloon-borne experiments underway include EBEX~\cite{Reichborn-Kjennerud10} and SPIDER~\cite{Filippini10}, and planned satellite missions include LiteBIRD~\cite{Matsumura14}). All of these experiments are fielding, or plan to field, large ($>10^{3}$ pixels/array) sub-Kelvin arrays of TES bolometers. Broadly, these experiments seek to map the microwave sky ($\sim$(30-300)~GHz) in both intensity and polarization on arcminute scales to several degree angular scales to improve our understanding of cosmology, the evolution of structure in the universe, galaxy clusters, and millimeter sources. High signal-to-noise polarization maps over degree angular scales in particular have the potential to provide unique sensitivity to the signatures of gravitational waves produced in the very early Universe, such as those produced in some inflationary models~\cite{Abazajian15}. The readout of these large sub-Kelvin detector arrays is complex, requiring novel superconducting electronics with thousands of components and interconnects. To reduce the cost and complexity of the readout and the thermal load on the cryogenics presented by warm electronics and cables, the detectors on these arrays must be multiplexed, with many groups of detectors read out and controlled by a much smaller number of wires and devices. In this proceedings, we describe the implementation and testing of the custom readout developed for the first \AdvACT high frequency (HF) array, which consists of $2024$ TESes operated at $\sim100$~mK. In \s{sec:advact} we describe the \AdvACT experiment. In \s{sec:multiplexing} we discuss multiplexing generally and then the specific multiplexing implementation chosen for \AdvACT in \s{sec:implementation}. In \s{sec:characterization}, we describe screening and characterization performed on readout components for the first \AdvACT array, and we conclude in \s{sec:summary} with a discussion of the implications of this work for future efforts.

\label{sec:summary} The readout components for the first \AdvACT array have been screened, and integration is complete. We have described the design and performance of the readout electronics for the array, and this readout has been used successfully to characterize the \AdvACT HF array, which will be deployed in mid-2016 and observe the CMB at \SI{150}{\giga\hertz} and \SI{230}{\giga\hertz}. The readout techniques developed for \AdvACT and described in this proceedings enables the readout of significantly larger arrays of TES bolometers than was previously possible using an MCE. While other multiplexing schemes employing alternate technologies are currently under development and could yield MUX factors an order of magnitude larger than the 64 MUX factor demonstrated here,~\cite{Day03,Irwin04,Kher16} current efforts that rely on low-frequency TDM to multiplex large arrays of TES bolometers stand to benefit substantially from larger MUX factors.

1607.06064

1607

1607.00312_arXiv.txt

{ At the Hamburger Sternwarte an effort was started in 2010 with the aim of digitizing its more than 45\,000 photographic plates and films stored in its plate archives. At the time of writing, more than 31\,000 plates have already been made available on the Internet for researchers, historians, as well as for the interested public. The digitization process and the Internet presentation of the plates and accompanying hand written material (plate envelopes, logbooks, observer notes) are presented here. To fully exploit the unique photometric and astrometric data, stored on the plates, further processing steps are required including registering the plate to celestial coordinates, masking of the plates, and a calibration of the photo-emulsion darkening curve. To demonstrate the correct functioning of these procedures, historical light curves of two bright BL Lac type active galactic nuclei are extracted. The resulting light curve of the blazar 1ES~1215+303 exhibits a large decrease in the magnitude from $14.25^{+0.07}_{-0.12}$ to $15.94^{+0.09}_{-0.13}$ in about 300~days, which proves the variability in the optical region. Furthermore, we compare the measured magnitudes for the quasar 3C~273 with contemporaneous measurements. }

The advent of photographic emulsion techniques to detect and store telescopic observations in the second half of the 19$^{th}$ century provided a challenge for contemporaneous data analysis techniques. Most of the astronomical objects captured by the plate emulsion were not available for scientific analysis and were essentially ignored, given the rather tedious techniques required to extract astrometric and photometric information on individual objects. There are two major problems with photographic plates: First, the information contained on the plates is not readily available in digital form, and second, the photographic emulsions have a finite lifetime and thus all information contained on any given plate will eventually be lost, since there are no backup plates. The latter issue is particularly relevant since the lifetime of emulsions is around one hundred years. In the current era of CCD imaging, powerful techniques have been developed by astronomers to extract digital data from CCD frames using the enormous processing power of present-day digital computers. Thus, it is a natural step to try to recover the unique information captured by historical observations using modern data analysis methods. In order to accomplish that goal, the information contained in the plate emulsion needs to be converted into some digital form with the help of a scanning process. Once the plate information is available in digital form, the aging process of the photographic emulsion, which inevitably takes place during the long term storage in plate archives, is practically terminated. If not cataloged and digitized, large quantities of astronomic photographic plates and films stored worldwide will be ultimately lost. Given the availability of suitable scanning and storage technology, projects to digitize the plates and to make the digital data available in the context of virtual observatories have been started in many countries. Leading in this context is the Harvard DASCH project \citep{dasch1}, many other observatories are now following the approach to digitize their plate collections, with many of them using commercial flatbed scanners of type Epson EXPRESSION 10000 XL. Hamburger Sternwarte, an observatory inaugurated in 1912 in Hamburg-Ber\-ge\-dorf, was designed for state-of-the-art research. At the time of its construction, all new telescopes were already equipped with cameras and photographic plates as detectors. For some telescopes objective prisms were available to accommodate the upcoming need for spectroscopic astrophysical research. During more than 80 years of observations in Hamburg, astronomical plates were exposed and most of them are still stored and available in the observatory's plate archives. These plates together with observers' notes and logbooks present a unique and when digitized also useful data base for modern long-term studies and demonstrate the development of astronomical and astrophysical research and observational techniques throughout the 20$^{th}$ century. In 2010 a pilot project was initiated to study the feasibility of a complete digitization and open access presentation of the data in the Internet. Already on November 1$^{st}$ 2011, the first 3\,700 scans of the very oldest plates were presented \footnote{\url{http://plate-archive.hs.uni-hamburg.de}\,(D.\,Groote)} and since then all scanned plates have been made publicly available immediately. In 2012, a digitization project was started by a consortium of Potsdam, Bamberg and Hamburg observatories, funded by the Deutsche Forschungsgemeinschaft (DFG). The goal of the project is to digitize all of the nearly 100~000 plates of the three observatories, to publish the data in the APPLAUSE database \citep[\textbf{A}rchive of \textbf{P}hoto\-graphic \textbf{PL}ates for \textbf{A}stro\-nomical \textbf{USE}, see e.g.][]{prague1, prague2}, and to make the data also available through the German Astrophysical Virtual Observatory (GAVO). At the time of writing, more than 31\,000 plates have already been digitized in Hamburg and are available for download in high resolution. The plates can be accessed through a web interface. Each photographic plate is presented on an individual web page along with the associated metadata and scans of observer notes and logbook entries, a low resolution image, and links to the high resolution scans. An easy-to-use archive search includes the emulsion types, telescope modes and filters as well as coordinate search and exposure times. In parallel to the digitization process, a data analysis framework has been developed that allows to retrieve astrometric and photometric data from the plates. To demonstrate the scientific value of the plate archives, we developed calibration procedures to extract otherwise unavailable historical information on bright BL~Lac type active galactic nuclei, which were ``unknown'' objects when the plates were taken. Yet these sources are notoriously variable and long-term variability studies are useful to characterize the stochastic nature of their variability. Here we will briefly describe the overall digitization project in Sect.~\ref{project}, and continue with the description of the automated astrometric and the photometric calibration in Sect.~\ref{analysis}. Our target selection and the results of the analysis of bright {BL Lac} type objects are given in Sect.~\ref{results} and discussed in Sect.~\ref{discussion}.\\

\label{discussion} The progress of the APPLAUSE project demonstrates that the content of the plate archives at Hamburger Sternwarte can be digitized, accessed through the Internet and utilized for scientific investigations. The digital data archive recovers a huge amount of material which can be used for scientific analysis. We estimate that a historical record of positions and magnitudes of about two billions of objects is now easily accessible to the scientific community. The photometric calibration of the data compares well with other, similar efforts and yields accuracies of typically 0.2-0.5 mag. In addition, a large number of spectra, normally taken with an objective prism, are also available on the plates of the Hamburger Sternwarte plate archives. While these spectra normally have relatively low resolution and can therefore be easily superseded by modern spectra, a number of historic spectra exist, for example of the NOVA GEM 1912. These spectra can now be extracted and used for modeling with state-of-the-art nova modeling tools. However, quite naturally, most of the work possible with the plate archives deals with light curves. \\ Our analysis of the plates stored in the plate archives of Hamburger Sternwarte shows that photographic plates can still be used effectively with modern data analysis, which we have shown to provide results consistent with previous findings. The research of highly variable objects like blazars or qua\-sars using these historical data allow glimpses into the past that would be otherwise impossible. Conclusions about the physical conditions can be drawn from fluctuations of the resulting long-time light curves, and photographic plates are the only medium to access recordings that date back so far. \\ Clearly, the analysis carried out, so far, covers only a fraction of the available potential; for example, only for about 50\% of the plates, the coordinates of the plate centers are listed in the metadata, while the rest is usually annotated by the name of the sky region observed, for example, ``Pleiades'', ``Kom. 1912d Westphal'' or similar, or is just empty. Normally the plate center of such plates can be reconstructed with high accuracy using the ``astronomy.net'' software. We expect that the full archive will be completed and available by mid 2017. \newpage

1607.00312

1607

1607.04707_arXiv.txt

We investigate the effect of strong magnetic fields on the adiabatic radial oscillations of hadronic stars. We describe magnetized hadronic matter within the framework of the relativistic nonlinear Walecka model and integrate the equations of relativistic radial oscillations to determine the fundamental pulsation mode. We consider that the magnetic field increases, in a density dependent way, from the surface, where it has a typical magnetar value of $10^{15}$ G, to the interior of the star where it can be as large as $3 \times 10^{18}$ G. We show that magnetic fields of the order of $10^{18}$ G at the stellar core produce a significant change in the frequency of neutron star pulsations with respect to unmagnetized objects. If radial pulsations are excited in magnetar flares, they can leave an imprint in the flare lightcurves and open a new window for the study of highly magnetized ultradense matter.

\label{sec:intro} Compact stars have a large number of pulsation modes that have been extensively studied since the seminal work of Chandrasekhar on radial oscillations \cite{APJ140:417:1964,PRL12:114:1964}. In general, these modes are very difficult to observe in the electromagnetic spectrum; therefore most efforts have concentrated on gravitational wave asteroseismology in order to characterise the frequency and damping times of the modes that emit gravitational radiation. In particular, various works focused on the oscillatory properties of pure hadronic stars, hybrid stars and strange quark stars trying to find signatures of the equation of state of high density neutron star matter (see \cite{AA325:217:1997,IJMPD07:29:1998,AA366:565:2001,APJ579:374:2002,PRD82:063006:2010,EL105:39001:2014} and references therein). More recently, compact star oscillations have attracted the attention in the context of Soft Gamma ray Repeaters (SGRs), which are persistent X-ray emitters that sporadically emit short bursts of soft $\gamma$-rays. In the quiescent state, SGRs have an X-ray luminosity of $\sim 10^{35}$ erg/s, while during the short $\gamma$-bursts they release up to $10^{42}$ erg/s in episodes of about 0.1 s. Exceptionally, some of them have emitted very energetic giant flares which commenced with brief $\gamma$-ray spikes of $\sim 0.2$ s, followed by tails lasting hundreds of seconds. Hard spectra (up to 1 MeV) were observed during the spike and the hard X-ray emission of the tail gradually faded modulated at the neutron star (NS) rotation period. The analysis of X-ray data of the tails of the giant flares of SGR 0526-66, SGR 1900+14 and SGR 1806-20 revealed the presence of quasi-periodic oscillations (QPOs) with frequencies ranging from $\sim$ 18 to 1840 Hz \cite{APJ628:L53:2005,APJ632:L111:2005,AA528:A45:2011}. There are also candidate QPOs at higher frequencies up to $\sim 4$ kHz in other bursts but with lower statistical significance \cite{ElMezeini2010}; in fact, according to a more recent analysis only one burst shows a marginally significant signal at a frequency of around 3706 Hz \cite{Huppenkothen2013}. Several characteristics of SGRs are usually explained in terms of the \textit{magnetar} model, assuming that the object is a neutron star with an unusually strong magnetic field ($B \sim 10^{15} $ G) \cite{Woods2006}. In particular, giant flares are associated to catastrophic rearrangements of the magnetic field. Such violent phenomena are expected to excite a variety of oscillation modes in the stellar crust and core. In fact, recent studies have accounted for magnetic coupling between the crust and the core, and associate QPOs to global magneto-elastic oscillations of highly magnetized neutron stars \cite{Levin2007,CerdaDuran2009,Colaiuda2012,Gabler2014}. There has also been interest in the possible excitation of low order $f$-modes because of their strong coupling to potentially detectable gravitational radiation \cite{Levin2011}. In the present paper we focus on radial oscillations of neutron stars permeated by ultra-strong magnetic fields. These modes might be relevant within the magnetar model because they could be excited during the violent events associated with gamma flares. Since they have higher frequencies than the already known QPOs, they cannot be directly linked to them at present. However, it is relevant to know all the variety of pulsation modes of strongly magnetized neutron stars because the number of observations is still small and new features could emerge in future flares' data. On the other hand, in the case of rotating objects we can expect some amount of gravitational radiation from even the lowest ($l = 0$) quasi-radial mode \cite{Stergioulas2003,Passamonti2006} making them potentially relevant for gravitational wave astronomy.

In this section we analyse the effect that a strong magnetic field could produce on the fundamental mode of the radial oscillations of hadronic stars. As mentioned before, we consider that the magnetic field decays with the density following the fast and slow profiles presented in Section \ref{B-density-depndence}. All the models for hadronic stars investigated in the present work have a maximum mass in agreement with the recent observation of the pulsars PSR J1614-2230 with $M = (1.97 \pm 0.04) M_{\odot}$ \cite{Demorest2010} and PSR J0348-0432 with $M = (2.01 \pm 0.04) M_{\odot}$ \cite{Antoniadis2013}. In Fig. \ref{fig1} we see that a magnetic field profile with $B_0 = 10^{17}$ G, produces very small changes on the oscillation period of the fundamental mode with respect to an unmagnetized star, for both slow and fast decays. This can be explained by the small effect that such magnetic field intensity has on the equation of state. In contrast, when $B_0 = 3.1 \times 10^{18}$ G is selected, there is a clear change in the oscillation period. As a function of the stellar mass, the curves fall below and to the right of the curves for weaker fields. Notice that for large mass objects the period changes because of the shift of the curves due to the increase of the maximum stellar mass. For smaller masses, the curves for $B_0 = 3.1 \times 10^{18}$ G are also significantly different with respect to the unmagnetized case; e.g. for a neutron star with $1.4 M_{\odot}$ the period is around $20 \%$ smaller, and the difference increases for less massive stars. For completeness we present also the oscillation period as a function of the gravitational redshift (see middle panel of Fig. \ref{fig1}) and as a function of the central mass-energy density (lower panel of Fig. \ref{fig1}). The effect of strong magnetic fields is more apparent in the frequency of the fundamental mode as can be seen in Fig. \ref{fig2}. The oscillation frequency for $B_0 = 10^{17}$ G is slightly above the one of an unmagnetized object of the same mass, and there is almost no difference between the fast and the slow decaying profiles of the magnetic field. However, if $B_0 = 3.1 \times 10^{18}$ G the oscillation frequencies are clearly larger than for an unmagnetized star of the same mass. For example, for a star with $1.4 M_{\odot}$ the frequency is around $20 \%$ larger and for a $1.7 M_{\odot}$ star it is around $10 \%$ larger. The difference between the fast and the slow decaying profiles of $B$ is very small. As stated before, purely radial modes do not emit gravitational waves and consequently they are essentially damped by the bulk viscosity, originated from the re-establishment of chemical equilibrium when a fluid element of the star is compressed and rarified during pulsations. Unfortunately, there is great uncertainty about the amount of viscosity inside neutron stars since it depends sensitively on the composition of matter which is uncertain beyond few times the nuclear saturation density \cite{Jones2001,Lindblom2002,Haensel2002,Chatterjee2007,Jha2010}. If the damping time due to viscous forces is long enough, radial pulsations in magnetars can leave an imprint in the microstucture of magnetar flare lightcurves opening a new window for the study of highly magnetized ultradense matter.

1607.04707

1607

1607.01314_arXiv.txt

We investigate the production of gravitational waves during preheating after inflation in the common case of field potentials that are asymmetric around the minimum. In particular, we study the impact of oscillons, comparatively long lived and spatially localized regions where a scalar field (e.g.\ the inflaton) oscillates with large amplitude. Contrary to a previous study, which considered a symmetric potential, we find that oscillons in asymmetric potentials associated with a phase transition can generate a pronounced peak in the spectrum of gravitational waves, that largely exceeds the linear preheating spectrum. We discuss the possible implications of this enhanced amplitude of gravitational waves. For instance, for low scale inflation models, the contribution from the oscillons can strongly enhance the observation prospects at current and future gravitational wave detectors.

1607.01314

1607

1607.06228_arXiv.txt

We summarize the results obtained from our suite of chemical evolution models for spiral disks, computed for different total masses and star formation efficiencies. Once the gas, stars and star formation radial distributions are reproduced, we analyze the Oxygen abundances radial profiles for gas and stars, in addition to stellar averaged ages and global metallicity. We examine scenarios for the potential origin of the apparent flattening of abundance gradients in the outskirts of disk galaxies, in particular the role of molecular gas formation prescriptions.

Chemical evolution models are the classical tool by which to interpret observed elemental abundances, and associated quantities such as gas and stellar surface densities, star formation histories, and the distribution of stellar ages. Elemental patterns carry the fingerprint of star formation timescales from their birth location, regardless of a star's present-day position. Chemical evolution codes solve a system of first order integro-differential equations, assuming an analytical star formation (SF) law, initial mass function (IMF), stellar lifetimes, and nucleosynthetic yields. In \cite{md05}, we calculated a grid of 440 theoretical galaxy models, (44 radial mass distributions, and 10 molecular gas and SF efficiencies between 0 and 1), calibrated on the Milky Way Galaxy (MWG). SF was assumed to occur in two steps: 1) molecular clouds forming from diffuse gas; 2) cloud-cloud collisions creating stars. Radial distributions for {\bf both} gas phases were derived, but the inferred predicted ratios of atomic to molecular gas, and SF rate ($SFR$), were found to be at variance with those observed. We are computing a new grid of models with updated stellar yields --\cite{molla15}--, gas infall rates --\cite{molla16a}--, and molecular gas formation efficiency --\cite{molla16b}. Our aim is to improve the predicted $H_{2}$ and $SFR$ profiles, while maintaining abundance radial gradients in agreement with those observed. We summarize our updated models and results in \S2 and our conclusions in \S3.

\begin{itemize} \item A grid of chemical evolution models with 16 dynamical masses in the range 10$^{10}$ to 10$^{13}\,M_{\odot}$ is calculated. A MWG-like model reproduces very well the observed radial distributions, as shown in \cite{molla15} and \cite{molla16b}. \item Prescriptions from \cite{bli} and \cite{fu10} for the formation of H$_{2}$ (BLI models) produce radial variations in O, $<Age>$, and $<Z/Z_{\odot}>$ which are stronger than STD models. Radial gradients are shown to be not invariant with radius. \item The H$_{2}$/HI relationship from \cite{bigiel08} is obtained for STD models, while BLI shows an unrealistically high dispersion. \end{itemize}

1607.06228

1607

1607.03127_arXiv.txt

We fit the rotation curves of isolated dwarf galaxies to directly measure the stellar mass-halo mass relation (\MstarM) over the mass range $5 \times 10^5 \simlt M_{*}/{\rm M}_\odot \simlt 10^{8}$. By accounting for cusp-core transformations due to stellar feedback, we find a monotonic relation with little scatter. Such monotonicity implies that abundance matching should yield a similar \MstarM\ if the cosmological model is correct. Using the `field galaxy' stellar mass function from the Sloan Digital Sky Survey (SDSS) and the halo mass function from the $\Lambda$ Cold Dark Matter Bolshoi simulation, we find remarkable agreement between the two. This holds down to $M_{200} \sim 5 \times 10^9$\,M$_\odot$, and to $M_{200} \sim 5 \times 10^8$\,M$_\odot$ if we assume a power law extrapolation of the SDSS stellar mass function below $M_* \sim 10^7$\,M$_\odot$. However, if instead of SDSS we use the stellar mass function of nearby galaxy groups, then the agreement is poor. This occurs because the group stellar mass function is shallower than that of the field below $M_* \sim 10^9$\,M$_\odot$, recovering the familiar `missing satellites' and `too big to fail' problems. Our result demonstrates that both problems are confined to group environments and must, therefore, owe to `galaxy formation physics' rather than exotic cosmology. Finally, we repeat our analysis for a $\Lambda$ Warm Dark Matter cosmology, finding that it fails at 68\% confidence for a thermal relic mass of $m_{\rm WDM} < 1.25$\,keV, and $m_{\rm WDM} < 2$\,keV if we use the power law extrapolation of SDSS. We conclude by making a number of predictions for future surveys based on these results.

\label{sec:intro} The standard $\Lambda$ Cold Dark Matter ($\Lambda$CDM) cosmological model gives an excellent description of the growth of structure in the Universe, matching the observed temperature fluctuations in the cosmic microwave background radiation \citep[e.g.][]{1992ApJ...396L...1S,2013arXiv1303.5076P}; the growth of large scale structure \citep[e.g.][]{2006Natur.440.1137S}; the clustering of galaxies \citep{2016MNRAS.455.4301C}; large scale weak lensing distortions \citep[e.g.][]{1991MNRAS.251..600B,2014MNRAS.441.2725F}; baryon acoustic oscillations \citep[e.g.][]{2003ApJ...594..665B,2005ApJ...633..560E,2013AJ....145...10D}; and the flux power spectrum of quasar absorption lines \citep[e.g.][]{1998ApJ...495...44C,2015arXiv151201981B}. However, over the past two decades there have been persistent tensions claimed on small scales inside galaxy groups and individual galaxies. These include: \begin{enumerate} \item {\it The `missing satellites' problem}: Pure dark matter cosmological simulations of structure formation predict that thousands of bound dark matter halos should orbit the Milky Way and Andromeda, yet only a few tens of visible satellites have been observed to date \citep[e.g.][]{1999ApJ...522...82K,1999ApJ...524L..19M,2012AJ....144....4M}. \item {\it The `cusp-core' problem}: These same simulations predict that the dark matter density distribution within galaxies should be self-similar and well fit at the ${\sim}10$\% level by the `NFW' profile (\citealt{1996ApJ...462..563N}): \begin{equation} \rho_{\rm NFW}(r) = \rho_0 \left(\frac{r}{r_s}\right)^{-1}\left(1 + \frac{r}{r_s}\right)^{-2} \label{eqn:rhoNFW} \end{equation} where the central density $\rho_0$ and scale length $r_s$ are given by: \begin{equation} \rho_0 = \rho_{\rm crit} \Delta c^3 g_c / 3 \,\,\,\, ; \,\,\,\, r_s = r_{200} / c \end{equation} \begin{equation} g_c = \frac{1}{{\rm log}\left(1+c\right)-\frac{c}{1+c}} \,\,\,\, ; \,\,\,\, r_{200} = \left[\frac{3}{4} M_{200} \frac{1}{\pi \Delta \rho_{\rm crit}}\right]^{1/3} \label{eqn:gcr200} \end{equation} $c$ is the dimensionless `concentration parameter'; $\Delta = 200$ is the over-density parameter; $\rho_{\rm crit}$ is the critical density of the Universe today; $r_{200}$ is the `virial' radius at which the mean enclosed density is $\Delta \times \rho_{\rm crit}$; and $M_{200}$ is the `virial' mass within $r_{200}$. For over two decades now, the rotation curves of small dwarf and low surface brightness galaxies have favoured a central constant density core over the `cuspy' \NFW\ profile described above \citep[e.g.][]{1994ApJ...427L...1F,1994Natur.370..629M,2002A&A...385..816D,2011MNRAS.414.3617K,2011AJ....142...24O,2013MNRAS.433.2314H}. \item {\it The `too big to fail' problem (TBTF)}: The central velocity dispersion of Local Group dwarfs appears to be too low to be consistent with the most massive subhalos in $\Lambda$CDM \citep{2006MNRAS.tmp..153R,2011MNRAS.415L..40B}. \end{enumerate} The above puzzles could be hinting at physics beyond $\Lambda$CDM, for example exotic inflation models \citep[e.g.][]{2002PhRvD..66d3003Z}, or exotic dark matter models \citep[e.g.][]{1994Natur.370..629M,2013MNRAS.430...81R,2014arXiv1412.1477E}. However, it is important to emphasise that all of these puzzles arise from a comparison between the observed Universe and a model $\Lambda$CDM universe entirely devoid of stars and gas (that we shall refer to from here on as `baryons'; e.g. see the discussion in \citealt{2014Natur.506..171P} and \citealt{2014JPhG...41f3101R}). Semi-analytic models make some attempt to improve on this by painting stars onto pure dark matter simulations \citep[e.g.][]{2006RPPh...69.3101B}. However, implicit in such analyses is an assumption that the distribution of dark matter is unaltered by the process of galaxy formation. It is becoming increasingly likely that this assumption is poor, especially within group environments and on the scale of tiny dwarf galaxies. \citet{1996MNRAS.283L..72N} were the first to suggest that dark matter could be collisionlessly heated by impulsive gas mass loss driven by supernova explosions. They found that, for reasonable initial conditions corresponding to isolated dwarf galaxies, the effect is small \citep[see also][]{2002MNRAS.333..299G}. However, \citet{2005MNRAS.356..107R} showed that the effect can be significant if star formation proceeds in repeated bursts, gradually grinding a dark matter cusp down to a core. There is mounting observational evidence for such bursty star formation \citep{2012ApJ...750...33L,2013MNRAS.429.3068T,2012ApJ...744...44W,2014MNRAS.441.2717K,2015MNRAS.450.3886M}, while the physics of such `cusp-core transformations' is now well-understood (\citealt{2012MNRAS.421.3464P,2015arXiv150207356P}, and for a review see \citealt{2014Natur.506..171P}). The latest numerical simulations that resolve the effect of individual supernovae explosions are substantially more predictive \citep[e.g.][hereafter R16a]{2015arXiv150202036O,2015MNRAS.454.2981C,2015arXiv150804143R}; these demonstrate that dark matter cores are an unavoidable prediction of $\Lambda$CDM (with baryons) for all low mass dwarf galaxies, so long as star formation proceeds for long enough\footnote{Two recent studies have claimed that dark matter cores do not form at any mass scale \citep{2016MNRAS.457.1931S,2016MNRAS.458.1559Z}. However, both of these used simulations with a `cooling floor' of $10^4$\,K, meaning that they are unable to resolve the clumpy interstellar medium. Resolving this is crucial for exciting cusp-core transformations, as explained in \citet{2012MNRAS.421.3464P}.}. However, there remains a debate in the literature over the efficiency of star formation in low mass halos. \citet{2014MNRAS.441.2986D}, \citet{2015MNRAS.454.2981C} and \citet{2015arXiv150703590T} find insufficient star formation to excite cusp-core transformations below $M_{200} \sim 10^{10}$\,M$_\odot$; \citet{2014ApJ...789L..17M} find that core formation proceeds in $M_{200} \sim 10^9$\,M$_\odot$ dwarfs; and R16a find that core formation proceeds `all the way down' to halo masses ${\sim}10^8$\,M$_\odot$. These differences owe in part to resolution. R16a have a typical spatial resolution of 4\,pc for their isolated dwarfs, with a stellar and dark matter particle mass resolution of ${\sim}250$\,M$_\odot$. This allows them to resolve the ${\simlt}500$\,pc size cores that form in their $M_{200} \simlt 10^9$\,M$_\odot$ dwarfs. Such small cores cannot be captured by the \citet{2014MNRAS.441.2986D} and \citet{2015arXiv150703590T} simulations that have a spatial resolution of ${\sim}80-100$\,pc. However, \citet{2015MNRAS.454.2981C} have a spatial resolution of ${\sim}30$\,pc for their $10^9$\,M$_\odot$ dwarf, yet they find that no significant dark matter core forms. This owes to a second key difference between these studies: the treatment of reionisation. In R16a, reionisation is not modelled and so star formation is allowed to proceed unhindered at very low halo mass. In all of the other studies, some model of reionisation heating is included. But the mass scale at which reionisation begins to suppress star formation, $M_{\rm reion}$, remains controversial. Some recent simulations favour a high $M_{\rm reion} \sim 10^{10}$\,M$_\odot$ \citep[e.g.][]{2015MNRAS.454.2981C,2015arXiv150703590T}, while others favour a much lower $M_{\rm reion} \sim 5 \times 10^8$\,M$_\odot$ \citep{2014ApJ...793...30G}, consistent with the assumption of no reionisation in R16a. Observationally, the continuous low star formation rate of nearby dwarf irregular galaxies (dIrrs) appears to favour a low $M_{\rm reion}$ \citep[][and see the discussion in R16a]{2009MNRAS.392L..45R,2012ApJ...748...88W}. We will discuss $M_{\rm reion}$ further in \S\ref{sec:reionisation}. Despite the differences in $M_{\rm reion}$, all of the above studies find that when dark matter cores do form, they are of size $\sim$ the projected half stellar mass radius ($R_{1/2}$). Such cores are dynamically important by construction because they alter the dark matter distribution precisely where we can hope to measure it using stellar kinematics (R16a). They also have important effects beyond just the internal structure of galaxies. Cored dwarfs are much more susceptible to tidal shocking and stripping on infall to a larger host galaxy \citep[e.g.][]{2006MNRAS.tmp..153R,2010MNRAS.406.1290P,2013ApJ...765...22B}. This aids in the morphological transformation of dwarfs from discs to spheroids \citep{2001ApJ...547L.123M,2012ApJ...751L..15L,2013ApJ...764L..29K}; and physically reshapes the dark matter halo mass function within groups \citep[][and see the discussion in R16a]{2010MNRAS.406.1290P,2012ApJ...761...71Z,2013ApJ...765...22B,2016arXiv160205957W}. Using simulations of isolated dwarfs at a spatial and mass resolution of ${\sim}4$\,pc and ${\sim}250$\,M$_\odot$, respectively, R16a derived a new `\coreNFW' fitting function that describes cusp-core transformations in $\Lambda$CDM over the mass range $10^8 \simlt M_{200}/{\rm M}_\odot \simlt 10^{10}$ (see equation \ref{eqn:coreNFW}). In \citet{2016arXiv160105821R} (hereafter R16b), we showed that this gives a remarkable match to the rotation curves of four isolated dwarf irregular galaxies, using just two free fitting parameters: $M_{200}$ and $c$ (that take on the same meaning as in equation \ref{eqn:rhoNFW} for the \NFW\ profile). In particular, using mock data, we demonstrated that if the data are good enough (i.e. if the dwarfs are not face-on; starbursting; and/or of uncertain distance) then we are able to successfully measure both $M_{200}$ and $c$ within our quoted uncertainties. In this paper, we apply the rotation curve fitting method described in R16b to 19 isolated dwarf irregulars (dIrrs) to measure the stellar mass-halo mass relation \MstarMrot\ over the stellar mass range $5 \times 10^5 \simlt M_{*}/{\rm M}_\odot \simlt 10^{8}$. We then compare this with the stellar mass-halo mass relation obtained from `abundance matching', \MstarMabund, to arrive at a comparatively clean test of our current cosmological model. This paper is organised as follows. In \S\ref{sec:newprobe}, we show how the comparison between \MstarMrot\ and \MstarMabund\ constitutes a rather clean cosmological probe at the edge of galaxy formation. In \S\ref{sec:data}, we describe our data compilation of rotation curves, stellar masses, and stellar mass functions. In \S\ref{sec:rotmethod}, we briefly review our rotation curve fitting method that is described and tested in detail in R16b. In \S\ref{sec:results}, we present the results from applying our rotation curve fitting method to 19 isolated dwarf irregular galaxies in the field (the individual fits and fitted parameters are reported in Table \ref{tab:data} and Appendix \ref{app:rotcurves}). In \S\ref{sec:discussion}, we discuss the implications of our results and their relation to previous works in the literature. Finally, in \S\ref{sec:conclusions} we present our conclusions.

\label{sec:conclusions} We have presented a clean probe of cosmology on small scales that follows from the comparison of \MstarMrot, measured from the rotation curves of isolated dwarf galaxies in the field, and \MstarMabund\ calculated from abundance matching (see \S\ref{sec:newprobe}). These should agree if the cosmological model is correct, but will diverge if the halo mass function is too shallow or steep on small scales. Our probe is comparatively clean since it relies only on the following theory ingredients: (i) a monotonic relation between stellar mass and halo mass; (ii) a predicted dark matter halo mass function; and (iii) a robust prediction of the internal dark matter distribution in dwarf irregular galaxies, for a given cosmological model. The first of these can be empirically tested using \MstarMrot; while (ii) and (iii) are readily obtained from state-of-the art numerical simulations (see \S\ref{sec:newprobe}). Our key results are as follows: \begin{itemize} \item We fit the rotation curves of a carefully selected sample of 19 isolated dIrr galaxies. Of these, five were found to be of too low inclination to be reliably inclination corrected (`inclination rogues'); another two (DDO 216 and NGC 1569) showed clear signs of disequilibrium (`disequilibrium rogues'); while one (DDO 101) had a very large distance uncertainty (`distance rogues'). For the remaining 11 dIrrs, we found that an \NFW\ dark matter halo profile is ruled out at $>99$\% confidence, reaffirming the well known `cusp-core' problem. By contrast, the \coreNFW\ profile from R16a -- that accounts for cusp-core transformations due to stellar feedback -- gives an excellent fit in all cases, without introducing any more free parameters than the \NFW\ form. \item Although we required the \coreNFW\ profile to obtain a good fit to the rotation curve shape, we showed that the implied dark matter halo mass $M_{200}$ was not sensitive to the form of the dark matter density profile within $r \simlt R_{1/2}$. For this reason, we were able to robustly measure the stellar mass-halo mass relation \MstarMrot\ over the mass range $5 \times 10^5 \simlt M_{*}/{\rm M}_\odot \simlt 10^{8}$, finding a monotonic relation with little scatter. \item Such monotonicity implies that abundance matching should yield a \MstarMabund\ relation that matches \MstarMrot, if the cosmological model is correct. Using the `field galaxy' stellar mass function from the Sloan Digital Sky Survey (SDSS) and the halo mass function from the $\Lambda$CDM Bolshoi simulation, we found remarkable agreement between the two. This held down to $M_{200} \sim 5 \times 10^9$\,M$_\odot$, and to $M_{200} \sim 5 \times 10^8$\,M$_\odot$ if we assumed a power law extrapolation of the SDSS stellar mass function below $M_* \sim 10^7$\,M$_\odot$. \item The good agreement between \MstarMrot\ and \MstarMabund\ means that there is no `missing satellites' or TBTF problem for our sample of isolated dIrrs down to at least $M_{200} \sim 5 \times 10^9$\,M$_\odot$. This is lower than the mass scale at which the `missing satellites' and TBTF problems manifest in the Local Group, $M_{\rm TBTF} \sim 10^{10}$\,M$_\odot$ \citep[e.g.][]{2006MNRAS.tmp..153R,2011MNRAS.415L..40B,2014MNRAS.440.3511T}. This suggests that both problems depend on {\it environment} and therefore owe to `galaxy formation physics' rather than exotic cosmology. \item Compiling stellar mass functions from the literature, we showed that the group stellar mass function is substantially shallower than the field below $M_* \sim 10^9$\,M$_\odot$. We argued that this likely owes to ram pressure stripping on group infall. This induces a significant scatter in $M_*$ for a given pre-infall $M_{200}$ causing classical abundance matching to fail. \item We considered how well a $\Lambda$ Warm Dark Matter ($\Lambda$WDM) cosmology can fit \MstarMrot. Repeating our abundance matching using the SDSS field stellar mass function, we showed that $\Lambda$WDM fails at 68\% confidence for a thermal relic mass of $m_{\rm WDM} < 1.25$\,keV, and $m_{\rm WDM} < 2$\,keV if we used the power law extrapolation of the SDSS stellar mass function. \item If $\Lambda$CDM is correct, we predict that the stellar mass function of galaxies should continue as an unbroken power law with slope $\alpha \sim 1.6$, at least over the mass range $10^5 < M_*/{\rm M}_\odot < 10^7$. There should be ${\sim}2000$ galaxies like Leo T in a typical 10\,Mpc$^3$ volume, with halo mass $5 \times 10^8 < M_{200}/{\rm M}_\odot < 10^9$; stellar mass ${\sim}2\times 10^5 < M_*/{\rm M}_\odot < 6 \times 10^5$; and HI gas mass ${\sim}3 \times 10^5 < M_{\rm HI}/{\rm M}_\odot < 3 \times 10^6$. Below this mass scale, we may see the first signs of star formation suppression due to reionisation. \end{itemize}

1607.03127

1607

1607.03916_arXiv.txt

{We report on a new mechanism that leads to the generation of primordial chiral gravitational waves, and hence, the violation of the parity symmetry in the Universe. We show that nonperturbative production of fermions with a definite helicity is accompanied by the generation of chiral gravitational waves. This is a generic and model-independent phenomenon that can occur during inflation, reheating and radiation eras, and can leave imprints in the cosmic microwave background polarization and may be observed in future ground- and space-based interferometers. We also discuss a specific model where chiral gravitational waves are generated via the production of light chiral fermions during pseudoscalar inflation. }

\label{sec:introduction} Gravitational waves (GW) are an invaluable probe for studying the early Universe as well as various astrophysical events. Recently, there has been a tremendous effort to detect them using both direct and indirect methods. Primordial gravitational waves, i.e., GW of cosmological origin, leave imprints on the polarization of the cosmic microwave background (CMB). There are several ongoing and upcoming experiments dedicated to observe the CMB polarization signal, such as Keck Array \cite{Array:2015xqh}, BICEP3 \cite{Ahmed:2014ixy}, PolarBEAR \cite{Ade:2014afa}, SPTpol \cite{Keisler:2015hfa} and ACTpol \cite{Niemack:2010wz}. In addition, the recent detection of GW from the astrophysical events GW150914 and GW151226 by LIGO-Virgo collaboration \cite{Abbott:2016blz,Abbott:2016nmj} has inspired the interest in the direct detection of the cosmological GW in ground and space-based interferometers. In fact, primordial GW will open a new observational window into the very first moments of the Universe. One of the important questions about the Universe is whether parity ($\cal P$) is respected or broken on macroscopic scales\footnote{Parity violation has a paramount importance in explaining the baryon asymmetry of the Universe. A possible connection between various baryogenesis mechanisms and macroscopic violation of $\cal P$ have been considered in Refs.~\cite{Joyce:1997uy,Brustein:1998du,Giovannini:1997eg,Giovannini:1999wv,Giovannini:1999by,Bamba:2006km,Bamba:2007hf,Fujita:2016igl,Domcke:2016bkh,Ben-Dayan:2016iks,Adshead:2016iae,Kamada:2016eeb,Anber:2015yca,Alexander:2016moy,Alexander:2004us,Adshead:2015jza,Long:2013tha,Sabancilar:2013raa}. }. Observation of chiral configurations of electromagnetic or gravitational fields on macroscopic scales will be a strong evidence for $\cal P$ violation. Chiral (circularly polarized) GW of primordial origin can give rise to non-zero TB and EB cross correlators in the CMB \cite{Gluscevic:2010vv}. Furthermore, they can also be directly detected in an array of ground and space-based interferometers \cite{Hayama:2016kmv,Aasi:2014erp,Cook:2011hg}. Only two mechanisms that can generate primordial chiral GW in the early Universe have been proposed in the literature. The first mechanism involves the coupling of a time varying scalar field to the gravitational topological term, i.e., $\Delta {\cal L} \propto \varphi R_{\mu\nu\rho\sigma} {\tilde R}^{\mu\nu\rho\sigma}$ \cite{Choi:1999zy,Lue:1998mq,Lyth:2005jf}. The other mechanism that produces chiral GW relies on the generation of helical gauge field configurations \cite{Garretson:1992vt,Anber:2006xt,Caprini:2014mja} via the term $\Delta {\cal L} \propto \varphi F_{\mu\nu} {\tilde F}^{\mu\nu}$ \cite{Sorbo:2011rz,Anber:2012du}. These helical gauge fields, in turn, contribute to the anisotropic stress tensor and source chiral GW. In this work, we report on a new model-independent mechanism that generates primordial chiral GW. We show that production of light chiral fermions in a time varying background is accompanied by chiral GW. Light chiral fermions with mass much smaller than the Hubble scale, $m/H \ll 1$, have definite helicity since the helicity flip process is negligible as it is suppressed by $(m/H)^2$. Production of fermions with a definite helicity leads to an asymmetry between the two components of the energy momentum tensor projected along the helicity eigenbasis. Hence, an imbalance between left and right helicities of GW will be created. Therefore, any mechanism that creates an asymmetry between light left and right-chiral fermions will also lead to chiral GW. This paper is organized as follows. In Sec.~\ref{sec:notation}, we start by developing the necessary formalism to study the generation of gravitational waves via the production of left-chiral Weyl fermions in the Fridmann-Robertson-Walker (FRW) background. Then, we compute the two point functions of the tensor perturbations in the helicity eigenbasis. At the end of Sec.~\ref{sec:notation}, we show that there is an imbalance between the correlation functions of the left and right-helical components of GW. Our main results are given by \eref{h correlator final} and \eref{right h correlator} for left and right-chiral fermions respectively. In Sec.~\ref{sec:inflation}, we give an explicit example where chiral GW are generated from the production of chiral fermions during pseudoscalar inflation. We conclude with a brief summary and discussion of our results in Sec.~\ref{sec:discussion}.

\label{sec:discussion} In this work, we reported on a new mechanism to generate chiral gravitational waves from the imbalance between left and right-handed fermions. This imbalance breaks $\cal P$ and leads to an enhancement of a certain helicity mode of gravitational waves. In particular, we showed that the nonperturbative production of chiral Weyl fermions in a time varying background is accompanied by tensor perturbations of preferred helicity, see \eref{h correlator final} and \eref{right h correlator} for left and right-chiral fermions, respectively. This mechanism can be generalized to any process that creates an asymmetry between fermions of different helicities. We also studied the generation of chiral GW from the production of fermions with a definite helicity in a model of pseudoscalar inflation \cite{Adshead:2015kza}. We calculated the power spectrum of the chiral components of the gravitational waves produced in this model. Chiral gravitational waves can be detected either indirectly using $\cal P$-odd TB and EB CMB correlators, which would otherwise be zero in the absence of $\cal P$ breaking, or using ground and space-based interferometers with polarization capabilities. In particular, we found that the amplitude of the difference between the chiral components of the superhorizon GW modes is not in reach within the current CMB polarization experiments. However, the situation changes if a large number of chiral fermions ${\cal N}$ is produced during inflation. Provided that the production of a large number of fermions does not backreact on inflation, the handedness of the tensor mode is \begin{eqnarray} \Delta \chi\cong {\cal N}\frac{H^2}{m_p^2}\,. \end{eqnarray} Thus, a large number of fermions\footnote{Large number of fermions can be realized in models that invoke large number of sectors, for instance, see the recent discussion of the so called ${\cal N}$Naturalness \cite{nnaturalness}.}, ${\cal N} \approx 10^{8}$ for $H\approx 10^{13}$ GeV, is needed in order to result in a detectable signal in current CMB polarization experiments. Also, we found that the generated chiral GW are almost scale invariant with a slight red tilt. It remains to be studied whether subhorizon modes of these chiral GW can be directly detected in ground and space-based interferometers. Particle production during inflation and radiation dominated eras can leave features in the primordial GW spectrum. Recently, there has been an interest in the study of models that can lead to the generation of gravitational waves that might be detected in future ground and space-based interferometers \cite{Figueroa:2013vif,Figueroa:2016ojl,Cook:2011hg}. However, most of these studies have focused on the magnitude of the GW rather than their polarization. It will be interesting to study the possibility of detecting the polarization of the primordial GW in these experiments. Applications of the mechanism we reported on in this work to various processes in the early Universe as well as the possibility of detecting chiral GW in future experiments are under our current investigation and will appear elsewhere.

1607.03916

1607

1607.04641_arXiv.txt

We present an X-ray and radio study of the famous `El Gordo', a massive and distant ($z=0.87$) galaxy cluster. In the deep (340~$\rm{ks}$) \chandra\ observation, the cluster appears with an elongated and cometary morphology, a sign of its current merging state. The GMRT radio observations at 610~$\rm{MHz}$ confirm the presence of a radio halo which remarkably overlaps the X-ray cluster emission and connects a couple of radio relics. We detect a strong shock ($\mathcal{M}\gtrsim3$) in the NW periphery of the cluster, co-spatially located with the radio relic. This is the most distant ($z=0.87$) and one of the strongest shock detected in a galaxy cluster. This work supports the relic--shock connection and allows to investigate the origin of these radio sources in a uncommon regime of $\mathcal{M}\gtrsim3$. For this particular case we found that shock acceleration from the thermal pool is still a viable possibility.

Galaxy clusters are the largest virialized structures in the Universe and form via aggregation of less massive systems \citep[\eg][]{press74}. During merger events, the intra-cluster medium (ICM) is heated by shocks and is believed to become turbulent. Part of the energy involved in these processes is converted into non-thermal phenomena that exhibit themselves in the radio band as halo and relic emissions \citep[\eg][for a review]{brunetti14rev}. Both radio sources are diffuse cluster-scale sources with steep spectra\footnote{$S_\nu \propto \nu^{-\alpha}$, with $\alpha$ spectral index.} ($\alpha\gtrsim1$). Radio halos are generally morphologically connected with the X-ray emission of the hosting cluster, whereas radio relics are elongated, polarized and found in cluster peripheries \citep[\eg][for an observational overview]{feretti12rev}. In particular, radio relics are believed to form at the gigantic shocks that are generated in major mergers, where cosmic ray electrons (CRe) are (re)accelerated \citep[see][for reviews]{bruggen12rev, brunetti14rev}. This scenario is supported by the arc-shaped morphologies of relics, their high level of polarization and by the fact that an increasing number of shocks have been detected at the location of radio relics \citep[\eg][]{akamatsu13systematic, bourdin13, shimwell15, eckert16, botteon16}. The main difficulty in the understanding of the origin of radio relics resides in the low Mach number ($\mathcal{M}\lesssim 3-4$) associated with merger shocks. The acceleration efficiency at these weak shocks is indeed expected to be small and in several cases it is in tension with the observational requirements \citep[\eg][]{markevitch05, macario11, kang12, pinzke13, kang14, vanweeren16toothbrush, botteon16}. \\ \indent ACT-CL J0102--4915 is the most massive cluster detected in the far Universe, at a redshift of $z=0.87$ \citep{menanteau12}. For its extraordinary mass of $M_{500}\sim8.8\times10^{14}$~$\rm{M_\odot}$ \citep{planck14xxix}, it is also known with the nickname of `El Gordo'. The cluster was firstly discovered by its strong Sunyaev-Zel'dovich (SZ) signal \citep{marriage11} and later confirmed through optical and X-ray observations. The system is in a complex merger state, as revealed by the double peaked galaxy distribution and the elongated morphology of its hot ($kT\sim15$ $\rm{keV}$) ICM \citep{menanteau10, menanteau12}. In the radio band, a tenuous halo and a double relic system at the cluster NW and SE X-ray boundaries were discovered \citep{lindner14}. \\ \indent In this paper we report the discovery of a strong shock associated with a radio relic in `El Gordo' cluster. In particular, our joint \chandra\ and \gmrt\ (GMRT) analysis provides interesting insights about the origin of the relic. Throughout the paper, we assume a concordance $\Lambda$CDM cosmology with $H_0 = 70$~$\rm{km\,s^{-1}\,Mpc^{-1}}$, $\Omega_{m}=0.3$ and $\Omega_{\Lambda}=0.7$, in which $1'' = 7.713$~$\rm{kpc}$ at the cluster redshift ($z=0.87$). Reported uncertainties are 68\%, unless stated otherwise.

We presented an X-ray/radio study of the famous `El Gordo' cluster located at $z=0.87$ focusing on the non-thermal activity in the cluster. \\ \indent Our GMRT radio observations at 610 and 327~$\rm{MHz}$ confirmed the presence of a halo and a system of double relics. These represent the most distant diffuse radio sources detected in a galaxy cluster so far. The halo is quite elongated in the NW-SE, \ie\ in the merger direction, and remarkably follows the ICM emission of the northern X-ray tail. The two relics are found at the boundaries of the X-ray emission. We focused on the NW relic which has a synchrotron spectral index $\alpha=1.37\pm0.20$ between 610 and 327~$\rm{MHz}$. \\ \indent The deep \chandra\ observations (340~$\rm{ks}$) allowed us to discover a shock at the position of the NW relic. The SB profile taken is this region abruptly drops at the relic location. The density compression factor $\mathcal{C} \gtrsim 3$ and the high downstream temperature provide the indication of a strong shock ($\mathcal{M} \gtrsim 3$) in the ICM. This is one of the three strongest shocks detected in galaxy clusters and the most distant ($z=0.87$) observed so far. \\ \indent The detection of a shock co-spatially located with a relic strongly supports the relic--shock connection. The NW shock in `El Gordo' cluster allows to study particle acceleration in a rare regime of strong shock. We found that DSA of thermal electrons is consistent with measured synchrotron spectrum. Nonetheless, only shocks with $\mathcal{M} > 3.5$ appear energetically viable while for weaker shocks re-acceleration models would be preferred. \\ \indent The presence of relativistic particles emitting a bright synchrotron relic at $z=0.87$ makes `El Gordo' a suitable cluster candidate to search for IC emission from the relic. From the X-ray spectral analysis we obtained possible hints for IC emission from the relic, however we could not firmly conclude the presence of IC excess and conservatively we derived only lower limits to the downstream magnetic field that have been used to improve constraints on particle acceleration. However, we also found hints of an excess in the 0.5-2~$\rm{keV}$ SB profile across the relic region. The combination of a possible IC excess in the spectral analysis with the hints of excess in the SB is tantalizing and certainly deserves deeper \chandra\ observations.

1607.04641

1607

1607.01587_arXiv.txt

We perform a linear stability analysis of magnetized rotating cylindrical jet flows in the approximation of zero thermal pressure. We focus our analysis on the effect of rotation on the current driven mode and on the unstable modes introduced by rotation. We find that rotation has a stabilizing effect on the current driven mode only for rotation velocities of the order of the Alfv\'en velocity. Rotation introduces also a new unstable centrifugal buoyancy mode and the ``cold'' magnetorotational instability. The first mode is analogous to the Parker instability with the centrifugal force playing the role of effective gravity. The magnetorotational instability can be present, but only in a very limited region of the parameter space and is never dominant. The current driven mode is characterized by large wavelenghts and is dominant at small values of the rotational velocity, while the buoyancy mode becomes dominant as rotation is increased and is characterized by small wavelenghts.

An important step for understanding the dynamics and phenomenology of astrophysical jets is the study of their instabilities. Instabilities have a substantial importance, on one hand for the formation and evolution of various observed structures and, on the other hand, for dissipating part of the jet energy and leading to the observed radiation. There are several possible sources of instabilities, like the velocity shear between the jet and the ambient medium, which drives the Kelvin-Helmholtz instability, the current flowing along magnetic field lines, which drives the current driven instability (CDI) and rotation that can drive several kinds of instabilities. Since the most promising models for the acceleration and collimation of jets involve the presence of a magnetic field with footpoints anchored to a rotating object (an accretion disk or a spinning star or black hole), the presence of a toroidal field component and of rotation seems to be a natural consequence and both CDI and rotation driven instabilities may play an important role in the jet propagation. CDI have been, for example, suggested as being responsible for the conversion from Poynting to kinetic energy flux in the first phases of jet propagation \citep{Sikora05}. KHI have been extensively studied in several different configurations both in the Newtonian \citep[see e.g.][]{Bodo89, Birkinshaw91, Hardee92, Bodo96, Hardee06, Kim15} and relativistic \citep[see e.g.][]{Ferrari78, Hardee79, Urpin02, Perucho04, Perucho10, Hardee07} cases. Similarly, CDI have been widely studied in the Newtonian limit \citep[see e.g.][]{Appl92, Appl96, Begelman98, Appl00, Baty02, Bonanno11, Bonanno11a}, while the analysis of the relativistic case has been more limited, most of the studies have considered the force-free condition \citep{Pariev94, Pariev96, Lyubarski99, Tomimatsu01, Narayan09} and only \citet{Bodo13} studied the full MHD case. The study of the effects of rotation have been mainly focused on the accretion disk problems, where the main instability considered is the magnetorotational instability \citep{Balbus92}, however, the combination of magnetic field and rotation can give rise to several other instabilities \citep[see e.g.][]{Kim00, Hanasz00, Keppens02, Varniere02, Huang03,Pessah05, Bonanno06, Bonanno07, Fu11} and the interplay between the different modes can become quite complex. Our goal is to study these rotation-induced instabilities in the context of magnetized jets. In \citet{Bodo13} (hereinafter Paper I) we studied the interplay between KHI and CDI in a relativistic non-rotating cold jet configuration, characterized by a current distribution concentrated inside the jet and closing at large distances. In this paper, we introduce the effects of rotation which, however, makes the analysis of the unstable modes much more intricate. Therefore, before tackling the full relativistic case, in this paper we limit ourselves to a newtonian analysis, neglecting again thermal pressure compared with the magnetic one. Moreover, since most of the unstable modes that we will consider are concentrated inside the jet radius and therefore the effect of the jet velocity would be to simply Doppler shift their frequencies, in this first step we ignored also the presence of the longitudinal velocity component. The main focus of this paper will then be on the effect of rotation on CDI and on the new modes of instability introduced by rotation. The effect of rotation on CDI was considered by \citet{Carey09} who analyzed a rigidly rotating jet and found a stabilizing effect for rotation periods shorter than a few Alfv\'en times. An analysis of the unstable modes introduced by rotation in a configuration and parameter range similar to ours has been performed by \citet{Kim00}. They discuss these modes in the cold plasma limit, however, their study is mainly local, whereas we focus more on global analysis of these instabilities, besides they do not discuss the CDI. Another related works are by \citet{Pessah05} and \citet{Huang03}, who examined the instabilities of axisymmetric perturbations in the presence of rotation and superthermal magnetic fields. The treatment of the first paper is again local and mostly focuses on an equilibrium configuration typical of accretion discs, while the second one analyses the stability a rotating cylindrical plasma Dean flow with only axial field both with local and global approach. The plan of the paper is the following: in the next section we present the equilibrium configuration, in Section \ref{sec:linequations} we derive the linearized equations, in Section \ref{modeclassification} we discuss the WKBJ local dispersion relation and present energetic considerations based on the Frieman-Rotenberg approach \citep{Frieman60} . The local dispersion relation and the energetic considerations will be useful in understanding the nature of the unstable modes that will be discussed in Section \ref{results}, where we present our results on global modes. Finally in the last Section \ref{summary} we summarize our findings.

\label{summary} \begin{figure} \centering \includegraphics[width=15cm]{fig19.eps} % \caption{\small The dominant instability types for $ka=0.1$, in the different regions of the parameter plane $(\Omega_c, P_c)$. In the region marked by a green shading, the instability with the largest growth rate is the CDI, while in the region marked by a blue shading the instability with the largest growth rate is the centrifugal buoyancy instability.} \label{fig:summary} \end{figure} We have examined the stability properties of a rotating magnetized jet flow in the approximation of zero thermal pressure. Our study has focused on the effect of rotation on the CDI and on the new modes of instability introduced by rotation. In this spirit, as a first step, we did not consider the presence of a longitudinal flow, whose main effect on the modes concentrated inside the jet radius, as it is for most of the rotationally-induced modes studied in this paper, is only that of Doppler shifting the frequency. The instability behaviour depends, of course, on the chosen equilibrium configuration and our results can be considered representative of an equilibrium configuration characterized by a distribution of current concentrated in the jet, with the return current assumed to be mainly found at very large distances. Similar stability analyses of rotation-induced modes in magnetized flows in the limit of zero thermal pressure are presented in \citet{Kim00, Huang03} and \citet{Pessah05}, they however make use only of a local WKB approach and the last two papers consider only axisymmetric perturbations. Our local analysis (see Section \ref{sec:wkb}) generally agrees with the results of these papers in the parameter regimes they consider. However, for our specific jet configuration we found that the MRI in the cold limit can be present, but only in a very limited region of the parameter space and is never dominant. We extended the results of these papers to the global domain, where the WKB approach no longer holds, by solving a boundary value problem and revealed new properties of these modes that are summarized below: \\ 1. At small and large axial wavenumbers $k$, the growth rates, respectively, for the toroidal and poloidal buoyancy modes obtained from the global calculations (Figs. 6, 11 and 15-18) actually exhibit a dependence on $k$ similar to what is obtained by the corresponding local dispersion relations (see Fig. 2). \\ 2. At intermediate $k$, the behavior can be different from that predicted on the basis of the local dispersion relation, with the presence of stability gaps (for positive $m$) and merging of poloidal and toroidal modes (for negative $m$). \\ 3. The properties of the MRI in the cold plasma limit, studied in the above papers based on the local dispersion relation, qualitatively agrees with our global calculations. In the nonaxisymmetric case, for positive $m$, the MRI is present only in a limited range of wavenumbers, its growth rate and the width of the unstable range reach a maximum for $ m=2$ and then (for larger $m$) decrease (Fig. 18) consistently with \citet{Kim00}, however, the MRI is absent for $m \geq 4$ and for every negative value of $m$. In addition, we did not find the MRI for $m=0$ in both local and global cases, likely because of the different equilibrium adopted. In general, the MRI has always a growth rate smaller than that of the buoyancy modes. We have shown that two main kinds of instabilities -- CDI and buoyancy -- prevail in the considered jet flow. In Fig. \ref{fig:summary} we represent, in the parameter plane $(\Omega_c, P_c)$, with different shadings, the regions where each of them has the largest growth rate. The figure refers to non-axisymmetric modes with $k=0.1$. We can observe that the CDI is dominant at small rotation rates and that the boundary between the CDI and centrifugal buoyancy instability regions moves towards larger values of $\Omega_c$ as we decrease $P_c$. For $P_c > 10$, the CDI is stable for this value of the wavenumber and the only instability is the centrifugal buoyancy, which is, however, obviously stable at zero rotation. It is seen that the buoyancy instability occupies quite a large area in this parameter space in comparison with the CDI and hence should be important in jets with rotation. When we increase the wavenumber, the CDI tends to be stabilized and the centrifugal instability tends to become dominant everywhere in the parameter plane. Comparing now the growth rates of axisymmetric and non-axisymmetric centrifugal buoyancy modes, we see that at high wavenumbers, axisymmetric modes have a larger growth rate, which decreases monotonically with decreasing $k$. By contrast, non-axisymmetric modes have a growth rate that is almost independent from the wavenumber and, therefore, become dominant at low values of $k$. Summarizing, at low rotation rates, the non-axisymmetric CDI is the instability that grows fastest and has large wavelengths. Increasing the rotation rate, the prevailing instability becomes the centrifugal axisymmetric one, which operates at small wavelengths. These results are applicable to magnetically and rotationally dominated jets, since, increasing the importance of thermal pressure, centrifugally driven modes tend to be stabilized and other modes, like pressure driven modes \citep{Kersale00} may appear. At the same time, taking into account the shear of longitudinal velocity can give rise to unstable KH modes in the jet. This first step will be extended by introducing the effects of the longitudinal velocity also in the relativistic regime and these results will be presented in a following paper. The different behaviour in the explored parameter space may be important for understanding the nonlinear stages since distinct types of instability may evolve differently. This study is therefore an essential first step for the interpretation of the results of numerical simulations and for their comparison with astrophysical data.

1607.01587

1607

1607.03061_arXiv.txt

We present 12\,mm Mopra observations of the dense ($>$10$^3$\,cm$^{-3}$) molecular gas towards the north-east (NE) of the W28 supernova remnant (SNR). This cloud is spatially well-matched to the TeV gamma-ray source HESS\,J1801$-$233 and is known to be a SNR-molecular cloud interaction region. Shock-disruption is evident from broad \nhone spectral line-widths in regions towards the W28 SNR, while strong detections of spatially-extended \nhthree, NH$_3$(4,4) and NH$_3$(6,6) inversion emission towards the cloud strengthen the case for the existence of high temperatures within the cloud. Velocity dispersion measurements and NH$_3$(n,n)/(1,1) ratio maps, where n=2, 3, 4 and 6, indicate that the source of disruption is from the side of the cloud nearest to the W28 SNR, suggesting that it is the source of cloud-disruption. Towards part of the cloud, the ratio of ortho to para-NH$_3$ is observed to exceed 2, suggesting gas-phase NH$_3$ enrichment due to NH$_3$ liberation from dust grain mantles. The measured NH$_3$ abundance with respect to H$_2$ is $\sim (1.2 \pm 0.5)\times10^{-9}$, which is not high, as might be expected for a hot, dense molecular cloud enriched by sublimated grain-surface molecules. The results are suggestive of NH$_3$ sublimation and destruction in this molecular cloud, which is likely to be interacting with the W28 SNR shock.

W28 is a mature ($>10^{4}$\,yr, \citealt{kaspi1993}) mixed-morphology supernova remnant (SNR) and a prime example of a region of TeV (10$^{12}$\,eV) gamma-ray excess overlapping with molecular gas \citep{hess:w28}; one indicator of a hadronic production mechanism. W28 is estimated to be at a distance of 1.2-3.3\,kpc (e.g. \citealt{goudis1976,lozinskaya1981,motogi2010}) and has been detected from radio to gamma-ray energies (e.g. \citealt{dubner2000,rho2002,hess:w28,fermi:w28,agile:w28,Nakamura:2014}). Of particular interest are the molecular clouds north-east (NE) of W28. Towards here, molecular emission lines have broad profiles \citep{arikawa1999,torres2003,reach2005,nicholas2011a, Nicholas:2012} and the presence of many 1720\,MHz OH \citep{DeNoyer:1983,frail1994,claussen1997} and 44\,GHz CH$_3$OH \citep{Pihlstrom:2014} masers suggest that the W28 SNR shock is disrupting the clouds. Furthermore, observations targeting the DCO$^+$/HCO$^+$ molecules in the north of these clouds suggest the presence of elevated levels of ionisation consistent with the existence of a nearby source of 0.1-1\,GeV cosmic rays \citep{Vaupre:2014}. In an attempt to understand the disruption and dynamics of all the molecular clouds surrounding W28, \citet{nicholas2011a} conducted broad scale ($\sim 1.5^{\circ}$\,square) observations of the W28 field in a 12\,mm line survey with $\sim 2^{\prime}$ FWHM resolution. The dense interiors of the molecular clouds towards the W28 TeV gamma-ray sources, HESS\,J1801$-$233 and HESS\,J1800$-$240 (sub-regions A, B and C), were probed with \nh3 inversion transitions observed at 12\,mm with the Mopra radio telescope. Multiple dense clumps and cores spatially-consistent with both the CO-traced gas, and TeV gamma-ray sources were revealed. Also, the extent to which the W28 SNR has disrupted the dense core of the NE cloud at line of sight velocity $\sim 7$\,kms$^{-1}$ was shown. Strong \nhthree emission, and NH$_3$(6,6) emission suggested this region is warm and turbulent \citep{nicholas2011a}. Modelling of the dense gas in the NE cloud with the MOLLIE radiative transfer software \citep{Keto:1990} suggested that the inner dense cloud component has mass $>1300$\,\msun. Further observations toward the W28 field were conducted in a 7\,mm line survey \citep{Nicholas:2012}, which offered superior angular resolution ($\sim 1^{\prime}$ FWHM) relative to 12\,mm observations. The J=1-0 transition of the CS molecule and isotopologues, C$^{34}$S and $^{13}$CS, were used as an independent probe of the dense gas in the region. Simultaneously-observed SiO(1-0) emission exposed the sites of shocks and/or outflows. Both CS(1-0) and SiO(1-0) emission were detected towards the NE cloud, revealing sub-structure in the shocked cloud that lower sensitivity \nh3 observations did not resolve. Figure\,\ref{fig:overview} indicates regions which have been mapped in previous molecular emission mapping campaigns towards the W28 SNR field. The broad spectral profiles from all lines detected in the NE cloud indicate that a kinetic energy of $\sim 10^{48}$\,erg is contained within turbulent gas motions \citep{nicholas2011a,Nicholas:2012}, and it is possible that multiple gas components exist. Detailed \nh3 spectra from across the entire cloud core may help to accurately determine the cloud temperature and density gradients, thus providing better constraints on the total dense cloud mass. Such constraints are important for investigations of the cosmic ray density in a hadronic scenario for gamma-ray emission (e.g. \citealt{Maxted:2013a,Maxted:2013b}), because the measured gamma-ray flux is proportional to both the gas mass and cosmic ray density. To further probe the structure of the dense and disrupted gas towards the NE of W28, the Mopra radio telescope is used to create deeper 12\,mm NH$_3$ inversion transition maps. These observations provide better sensitivity than any previous large scale dense gas studies of the region.% \begin{figure} \includegraphics[width=\columnwidth]{W28_90cm.eps} \caption{An image of 90\,cm continuum emission (linear 0.06-0.55\,Jy\,beam$^{-1}$) observed with the VLA \citep{brogan2006}. TeV gamma-ray emission significance (3-6\,$\sigma$) contours (thick, black) and CO\,(2-1) emission contours (dashed, blue) are overlaid \citep{hess:w28,fukui2008}.The approximate boundaries of catalogued SNRs \citep{brogan2006,yusef2000} are displayed as black thin circles.% Regions from previous 12\,mm mapping campaigns (including HOPS data from \citealt{walsh2008}) are enclosed by a thin black dashed boundary, whereas thin, solid magenta boundaries indicated previous 7\,mm mapping regions. Thick dashed black boxes indicate the mapped regions in this work.} \label{fig:overview} \end{figure}

\label{sec:conc} We reported on deep mapping observations towards the shocked molecular cloud north-east of W28, with a focus on detecting multiple \nh3 inversion transitions. % The NE cloud has a remarkable spatial match with the gamma-ray source HESS J1801-233, so constraints on the mass distribution are important for hadronic gamma-ray production models of the region, while the observed chemistry serves as an observational constraint on CR ionisation and propagation. Spectral line observations are steps towards parameter constraints associated with the NE cloud of W28. These observations revealed that the dense component of the NE cloud is much more extended than previously reported. This is the case for all the detected inversion transitions. Towards the cloud, strong \nhthree, NH$_3$(4,4) and NH$_3$(6,6) emission suggest this is a region of high gas temperature. Furthermore, new evidence for shocked gas is provided by \nhone\,-\nhthree velocity dispersion maps that resolve a new \nh3 component on the W28 side of the NE cloud. % Gas parameter maps were derived from NH$_3$ emission via a method that assumes Local Thermodynamic Equilibrium (LTE) and were checked against non-LTE statistical equilibrium models. NH$_3$ column densities on the order of 10$^{14}$\,cm$^{-2}$ and temperatures in the range 35-60\,K were observed within the NE cloud of W28. The ortho-para-NH$_3$ ratio (OPR) was investigated, revealing a subregion with an elevated OPR ($>$2), characteristic of regions where NH$_3$ is liberated from dust-grain mantles. Comparing our measurements with a previously-published CS-derived mass estimate, no corresponding NH$_3$ abundance enhancement was observed ([\nh3]/[H$_2$]$\sim (1.2 \pm 0.5)\times10^{-9}$), possibly suggesting the existence of an NH$_3$ destruction mechanism. More detailed modelling of gas-phase NH$_3$ production and destruction may be required to investigate this result. Future work to improve the angular resolution and sensitivity of TeV gamma-ray images will allow a detailed comparison of the gamma-ray emission and cosmic ray target material (the gas), while considering the time-dependent effect cosmic ray propagation may also allow the analysis of features in the GeV to TeV gamma-ray spectrum towards SNRs (e.g. \citealp{Gabici:2007, Maxted:2012}).

1607.03061

1607

1607.03347_arXiv.txt

We computed the contribution of the Compton and Bremsstrahlung processes with a hidden light $U(1)_D$ neutral boson $\gamma_D$ to the white dwarf G117-B15A anomalous cooling rate, as well as the white dwarf luminosity functions (WDLF). We demonstrated that for a light mass of hidden photon ($m_{\gamma_D} \ll$ a few keV), compatible results are obtained for the recent Sloan Digital Sky Survey and the SuperCOSMOS Sky Survey observation, but the stringent limits would be imposed on the kinetic mixing $\epsilon$. We performed $\chi^2$-tests to acquire a quantitative assessment on the WDLF data in the context of our model, computed under the assumption of different kinetic mixing $\epsilon$, the age of the oldest computed stars $T_D$, and a constant star formation rate $\psi$. Then taken together, the WDLF analysis of 2$\sigma$ confidence interval $\epsilon = \left( 0.37^{+0.35}_{-0.37}\right) \times 10^{-14}$ is barely consistent with the cooling rate analysis at 2$\sigma$ regime $\epsilon = \left( 0.97^{+0.35}_{-0.37} \right) \times 10^{-14}$. The two approaches used here agree with each other in yielding an anomalous cooling rate of white dwarf in this luminosity range.

The addition of an extra $U(1)_D$ gauge symmetry, implying the existence of an exotic massive neutral gauge boson, is one of the much investigated extensions of the Standard Model (SM) \cite{SM}. This exotic gauge boson is called hidden(dark) photon and denoted by $\gamma_D$ in literatures \cite{Holdom,Nath,HeCalculation,He1,SB,Foot,nonFoot,Hooper1,Hooper2,figure,g-2,My,HZZp-LHC,TC3}, and the interaction between $\gamma_D$ with the Standard Model particles arise only through a kinetic mixing $\epsilon$ \cite{Holdom}. The extra $U(1)_D$ gauge field is predicted in some phenomenology theories \cite{Nath,HeCalculation,He1,SB,Foot,nonFoot,Hooper1,Hooper2,figure,g-2,My,HZZp-LHC,TC3}, in which $\gamma_D$ plays the role of a dark matter candidate. In these theories $\gamma_D$ is assumed to be a product in high energy processes. As a result we can directly or indirectly measure the trace of hidden photon in experiments and observations. In this paper, we describe the hidden photon production as well as the cooling anomaly in white dwarf. The original hint of a cooling anomaly came from the measurement \cite{OriginalG117}, of the rate of period change $\dot{P}$ for the 215s mode of G117-B15A \cite{G117,G1172}, which is significantly larger than the prediction of standard pulsation theory \cite{Isern1992}, see TABLE.\ref{table1}. According to previous researches \cite{OriginalG117,Isern1992,G1172,G117}, the relative rate of change of the white dwarf pulsation period $\dot{P}/P$ is essentially proportional to the temperature cooling rate $\dot{T}/T$, which is in accordance with the assumption that the white dwarf (ZZ ceti stars) is not yet crystallized \cite{Isern1992}. Hence if the anomalous cooling rate as implied by the excess rate of change of the pulsation period are indeed induceed by the hidden photons, one would get\cite{Isern1992,151208108} \begin{align} \label{LST} \frac{L_{\gamma_D}}{L_{st}} \simeq \frac{\dot{P}_{obs}}{\dot{P}_{th}}-1, \end{align} where $L_{st}$ is the standard luminosity of the pulsation white dwarf. This relation is available only when $L_{\gamma_D} \geq L_{st}$ from the results reported in the reference research \cite{OriginalG117}. After the original measurement of \cite{OriginalG117}, astrophysicists also observed other two pulsation white dwarfs in two decades. The most present results \cite{151208108} are shown in TABLE.\ref{table1}, in which the data of two DA white dwarfs G117-B15A \cite{G1172,G117}, R548 \cite{R548} and a DB white dwarf PG 1351+489 are presented \cite{PG1351}. As this table, the constraint from G117-B15A is stronger then others. \begin{table}[b!] \setlength{\arrayrulewidth}{0.95pt} \tabcolsep=4pt \renewcommand{\arraystretch}{0.2} \caption{\baselineskip 14pt Results of $\dot{P}$ for G117-B15A \cite{G117}, R548 \cite{R548}, PG 1351+489 \cite{PG1351}. This table is referenced from \cite{151208108}.} % \centering % \begin{tabular}{c c c c c } % \hline\hline % WD & class $\quad$ & $\dot{P}_{obs}[\hbox{s/s}]$ & $\quad$ $\dot{P}_{th}[\hbox{s/s}]$ & $\dot{P}_{obs} / \dot{P}_{th}-1$ \\ [1ex] \hline G117 - B15A & $\quad$ DA $\quad$ & $(4.19 \pm 0.73) \; \times \; 10^{-15}$ & $\quad$ $(1.25 \pm 0.09) \; \times \; 10^{-15}$ & $2.35 \pm 0.63$ \\ [1ex] R548 & $\quad$ DA $\quad$ & $(3.33 \pm 1.1) \; \times \; 10^{-15}$ & $\quad$ $(1.1 \pm 0.09) \; \times \; 10^{-15}$ & $2.03 \pm 1.03$ \\ [1ex] PG 1351+489 & $\quad$ DB $\quad$ & $(2.0 \pm 0.9) \; \times \; 10^{-13}$ & $\quad$ $(0.81 \pm 0.5) \; \times \; 10^{-13}$ & $1.47 \pm 1.89$ \\ [1ex] \hline\hline % \end{tabular} \label{table1} % \end{table} The first luminosity function was derived almost fifty years ago \cite{firstLF}, and during the long development period, it has been improved with significant amoumt of research works \cite{SSD1,SSD2,SSD3,SuperCOSMOS,Miller,Isern1,Iben,WDLFx,151208108,Wood1992,Salpeter1955,cutLF,OLF}. The recent availability of data are contributed by the Sloan Digital Sky Survey (SDSS) \cite{SSD1,SSD2,SSD3} and the SuperCOSMOS Sky Survey (SSS) \cite{SuperCOSMOS}, and those has noticeably improved the accuracy of the new luminosity functions. A recent analysis of the white dwarf luminosity functions (WDLF) is done by \cite{Miller} in which a unified WDLF is constructed by averaging the SDSS and SSS, and estimated the uncertainties by taking into account the differences between the WDLF at each magnitude bin. We use the \texttt{LPCODE} stellar evolution code\footnote{ \texttt{LPCODE} website: \url{fcaglp.fcaglp.unlp.edu.ar/evolgroup}} \cite{LPCODE29} for our stellar evolution computations. This code has been employed to study different problems related to the formation and evolution of white dwarfs \cite{LPCODE28,LPCODE29,LPCODE30,LPCODE0,LPCODEre}. In the following paragragh, we briefly outline the algorithm work of \texttt{LPCODE}, and more details are given by \cite{LPCODE29,14067712}. Radiative opacities are those of \cite{LPCODE31} while conductive opacities are cited by \cite{LPCODE32}, complemented at low temperatures (molecular opacities), which are produced by \cite{LPCODE33}. The equation of state for the high density regime is cited by \cite{LPCODE34}, while for the low density regime, an updated version of the equation of state of \cite{LPCODE35} is used. Neutrino cooling by pair, photo-, plasma, Bremsstrahlung production are included following the results of \cite{LPCODE36}, while plasma processes are also included \cite{LPCODE37}. White dwarf models computed with \texttt{LPCODE} also include detailed non-gray model atmospheres \cite{LPCODE38}. In addition, the effects of time dependent element diffusion during the white dwarf evolution are following by \cite{LPCODE39} for multicomponent gases. Finally, we have to mention that \texttt{LPCODE} has been tested against with other white dwarf evolutionary codes, and the discrepancies are lower than 2\%, see \cite{LPCODE40}. In Sec.$\;$2, we set up our model and also briefly discuss the mixing between the three neutral gauge bosons in the model as studied previously \cite{Nath,My}. In Sec.$\;$3, we present the calculations of hidden photon Compton scattering and Bremsstrahlung processes on white dwarf conditions. In Sec.$\;$4, we study that the effect contributed to hidden photon on white dwarf anomalous cooling rate and WDLF. We conclude our results in Sec.$\;$5.

In this paper we have derived an improved value of the kinetic mixing parameter $\epsilon$, assuming that the mass of the hidden photon $\gamma_D$ is smaller than the core temperature of white dwarfs ($\approx$ keV), and also assuming that the enhanced rate of cooling of the pulsating white dwarf is entirely due to the emission of hidden photons. In our calculations, we adopt the stellar evolution code \texttt{LPCODE} \cite{LPCODE29}, and ignore the density of hidden photons in white dwarf core, see Eq.~(\ref{Q}). Based on our assumptions, we found that the anomalous cooling rate of G117-B15A indicates the existence of an additional cooling mechanism in the pulsating white dwarf, consistent with kinetic mixing parameter $\epsilon = (0.97^{+0.18}_{-0.18}) \times 10^{-14}$, see Fig.~\ref{G117B15AFig}. In addition, we analyzed the contribution of hidden photons to the WDLF, see Fig.~\ref{LuminosityFunctionFig2}. We quantitatively weighted the agreement between theory and observations by means of $\chi^2$ fits, from which it follows that the kinetic mixing parameter $\epsilon = (0.37^{+0.35}_{-0.37}) \times 10^{-14}$ at a 95\% confidence level ($i.e. \approx 2\sigma$-like), and the best fit value is $\epsilon = 0.372 \times 10^{-14}$ which lies in the 1.64$\sigma$ confidence region. We compared the anomalous cooling rate and WDLF data, and found that the WDLF 2$\sigma$ confidence as shown in Fig.~\ref{LuminosityFunctionFig3} ($i.e. \; \epsilon \leq 0.72 \times 10^{-14}$) is compatible with the cooling rate 2$\sigma$ confidence in Fig.~\ref{G117B15AFig} ($i.e. \; \epsilon = 0.97^{+0.35}_{-0.37} \times 10^{-14}$). Both approaches agreed with each other in confirming the existence of an anomalous rate of cooling of white dwarf with $\epsilon \simeq (0.60 \sim 0.72) \times 10^{-14}$ in this luminosity range. Our results also indicate that hidden photons are dominant in white dwarf radiations within the range of $6.8 \leq M_{Bol} \leq 10.5$, see Fig.~\ref{LuminosityFunctionFig1}. It is important to emphasize that both methods are complementary and equally sensitive to the emission of hidden photons in white dwarf.

1607.03347

1607

1607.08749_arXiv.txt

The Large Binocular Telescope Interferometer uses a near-infrared camera to measure the optical path length variations between the two AO-corrected apertures and provide high-angular resolution observations for all its science channels (1.5-13~$\upmu$m). There is however a wavelength dependent component to the atmospheric turbulence, which can introduce optical path length errors when observing at a wavelength different from that of the fringe sensing camera. Water vapor in particular is highly dispersive and its effect must be taken into account for high-precision infrared interferometric observations as described previously for VLTI/MIDI or the Keck Interferometer Nuller. In this paper, we describe the new sensing approach that has been developed at the LBT to measure and monitor the optical path length fluctuations due to dry air and water vapor separately. After reviewing the current performance of the system for dry air seeing compensation, we present simultaneous H-, K-, and N-band observations that illustrate the feasibility of our feedforward approach to stabilize the path length fluctuations seen by the LBTI nuller.

\label{sec:intro} % Water vapor turbulence can limit the performance of high dynamic range interferometers that use phase-referenced modes, even though such turbulence is only a small contributor to the total seeing at visible and infrared wavelengths. Indeed, because water vapor is highly dispersive, random fluctuations in its differential column density above each aperture (or water vapor seeing) will create a chromatic component to the optical path difference (OPD) that is not properly tracked at wavelengths different from that of the fringe sensor. The impact of this effect on infrared interferometry has been addressed extensively in the literature, either in a general context\cite{Colavita:2004} or applied to specific instruments that include phase-referenced modes using K-band light such as VLTI/MIDI\cite{Meisner:2003,Matter:2010,Pott:2012,Muller:2014}, VLTI/GENIE\cite{Absil:2006}, or the Keck Interferometer Nuller\cite{Koresko:2003,Koresko:2006,Colavita:2010} (KIN). In this paper, we address the effect of precipitable water vapor (PWV) in the context of the Large Binocular Telescope Interferometer\cite{Hinz:2008} (LBTI), which uses a K-band fringe sensor to provide high-bandwidth path length compensation for all its scientific channels (1.5-13\,$\upmu$m) and in particular for the nulling interferometer that operates at 11\,$\upmu$m.

\label{sec:summary} Recent efforts in co-phasing the LBTI were focused on two main areas: (1) improving the performance of the system at K band where the fringe sensor operates and (2) compensating for atmospheric dispersion with the implementation of a new sensing mode. Thanks to an improved real-time control loop, a better vibration environment, and the routine use of OPD feedforward, the LBTI currently achieves a co-phasing stability of 280\,nm RMS at 1\.kHz. This performance is currently limited by high-frequency phase fluctuations occurring inside the LBTI cryostat and future work will focus on identifying and damping the source of these vibrations. Regarding atmospheric dispersion, a new mode was implemented to measure the phase at H and K band simultaneously (each using a different output of the interferometer). Using the information provided by the two wavelengths, it is possible to predict the phase variations at N band where the LBTI nuller operates. We have recently demonstrated the feasibility of this approach with simultaneous H-, K-, and N-band data. Future work will focus on using this signal to adjust the null setpoint in real time and, therefore, mitigate the phase variations in the nuller bandpass.

1607.08749

1607

1607.04567_arXiv.txt

\bicepthree\ is a small-aperture refracting cosmic microwave background (CMB) telescope designed to make sensitive polarization maps in pursuit of a potential \bmode\ signal from inflationary gravitational waves. It is the latest in the \bicep/\keckarray\ series of CMB experiments located at the South Pole, which has provided the most stringent constraints on inflation to date. For the 2016 observing season, \bicepthree\ was outfitted with a full suite of 2400 optically coupled detectors operating at 95 GHz. In these proceedings we report on the far field beam performance using calibration data taken during the 2015-2016 summer deployment season \emph{in situ} with a thermal chopped source. We generate high-fidelity per-detector beam maps, show the array-averaged beam profile, and characterize the differential beam response between co-located, orthogonally polarized detectors which contributes to the leading instrumental systematic in pair differencing experiments. We find that the levels of differential pointing, beamwidth, and ellipticity are similar to or lower than those measured for \biceptwo\ and \keckarray. The magnitude and distribution of \bicepthree 's differential beam mismatch -- and the level to which temperature-to-polarization leakage may be marginalized over or subtracted in analysis -- will inform the design of next-generation CMB experiments with many thousands of detectors.

\label{sec:intro} % A period of accelerated exponential expansion in the early Universe, while a radical extrapolation from understood physics, naturally solves the flatness, horizon, and monopole problems of standard cosmology\cite{planck_p15xx_15}. This inflationary period also explains the origin of structure by stretching quantum fluctuations to macroscopic scales. Inflation predicts perturbations to the metric in both scalars (density waves) and tensors (gravitational waves). Both types of perturbations affect the polarization of the cosmic microwave background (CMB) at last scattering\cite{polnarev85}: scalars can only generate a gradient-type (\emode) pattern, while tensors can also generate a curl-type (\bmode) pattern. Measurement of \bmode\ power in the CMB at degree angular scales in excess of the expectation from gravitational lensing would be direct evidence for an inflationary period, and the amplitude of the signal -- parametrized by $r$, the tensor-to-scalar ratio -- would indicate the energy scale of inflation. Since 2006, the \bicep/\keckarray\ CMB polarization experiments have been mapping $\sim 1\%$ of low-foreground sky from the Amundsen-Scott South Pole Station. The telescopes are optimized to detect fluctuations at degree angular scales, and therefore use a compact, on-axis refracting optical system with the minimum aperture necessary to resolve the $\ell \sim 100$ peak in the inflationary gravitational wave \bmode\ spectrum. In 2014 we reported a detection of \bmode\ power at degree angular scales at 150 GHz by \biceptwo\cite{biceptwo14_I}, which was quickly confirmed by data from the \keckarray\cite{biceptwo15_V}. Multifrequency maps are required to separate Galactic foreground emission (thermal dust at high frequencies and synchrotron at low frequencies) from the CMB signal. Using all available \planck\ and \wmap\ maps and the latest \keckarray\ maps at both 95 and 150 GHz, we find that the excess at 150 GHz is consistent with dust emission, and have constrained the tensor-to-scalar ratio to $r < 0.07$ at $95\%$ confidence\cite{biceptwo15_BKP_all,biceptwo16_VI}. To push this limit further -- or achieve a detection of primordial \bmode s -- we have deployed 220 GHz detectors in the \keckarray\ to measure dust at high precision, and have deployed \bicepthree\cite{wu15} at 95 GHz to probe deeply at a low-foreground frequency. For the 2016 observing season, \bicepthree\ was outfitted with a full complement of 20 detector tiles (2400 optically coupled detectors). The \bicep/\keckarray\ telescopes measure polarization by differencing co-located, orthogonally polarized detector pairs\cite{biceptwo14_II}. Any mismatch in beam shape between the two polarizations will leak some of the bright CMB temperature signal into polarization. Temperature-to-polarization leakage is the most prominent instrumental systematic effect which may cause a false \bmode\ signal and must be precisely characterized\cite{biceptwo15_IV}. We have developed an analysis technique called deprojection\cite{biceptwo15_III} to marginalize over any false polarization signal which would arise from the coupling of the CMB temperature sky to a second-order expansion of the measured differential beam; however, the contribution from higher-order beam mismatch remains. In the \biceptwo\ final results, using high-fidelity far field beam maps we constrained the undeprojected residual contribution to the final \bmode\ spectrum to an equivalent $r < 3.0 \times 10^{-3}$. In these proceedings, we present an analysis of the far field beam response of the \bicepthree\ detectors observing in the 2016 season. In February and March of 2016, we obtained far field beam maps of all \bicepthree\ detectors \emph{in situ} at the South Pole using a chopped thermal source in the far field of the telescope. These measurements serve several purposes: the array-averaged beam profile is used to smooth \healpix\ maps for simulations, the measured beam shapes offer a cross-check that the modes removed by deprojection correspond to actual beam mismatch and also allow for direct subtraction of the deprojection templates, and the beam maps are used in dedicated simulations to explicitly predict the residual \bmode\ power after deprojection. In Section~\ref{sect:optics} we review the optical design of \bicepthree. In Section~\ref{sect:ffbm} we describe the far field beam measurement setup and the new chopped source constructed in 2016. In Section~\ref{sect:beamparam} we present two-dimensional Gaussian fits to each detector and characterize beamwidth, ellipticity, differential pointing, differential beamwidth, and differential ellipticity across the entire array. In Section~\ref{sect:composites} we coadd all beam measurements to obtain high signal-to-noise composite maps of individual beams, and use these beams to calculate the array-averaged beam response. In Section~\ref{sect:deprojection} we discuss the measured beam mismatch in the context of deprojection and the residual beam systematics levels achieved by \biceptwo\ and \keckarray. Three companion papers -- a \bicepthree\ status update\cite{grayson16}, \bicepthree\ detector performance\cite{hui16}, and measurements of \keckarray\ detector polarization angles with a dielectric sheet calibrator\cite{bullock16} -- are also presented at this conference.

\label{sect:conclusions} In these proceedings we have presented a preliminary analysis of \bicepthree's far field beam performance for the 2016 observing season. Measurements were made during the 2015-2016 deployment season using a new chopped thermal microwave source with a 24'' aperture. We find that \bicepthree\ has a median beamwidth of $0.167^{\circ}$ (23.6 arcminutes) and present the array-averaged $B(\ell)$ profile. When detector pairs are fit to 2D elliptical Gaussians, we find that differential pointing, beamwidth, and ellipticity are similar to or lower than those measured for \biceptwo\ and \keckarray. The high-fidelity per-detector composite beam maps we have constructed from 77 beam mapping schedules will allow for explicit simulation of the undeprojected residual temperature-to-polarization leakage expected in the final polarization maps produced by \bicepthree.

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