adsabs/WIESP2022-NER huggingface dataset:
Paper: 2019A&A...621A.121W (fulltext)
the exact influence of radiation pressure on nitrogen atoms would require a dynamical model, such a low value for β indicates that it should have a negligible effect. Despite these very weak effects from radiation pressure, we do not find an overabundance of nitrogen in the β Pictoris disc. Other papers have also shown no clear deficiency in radiation-sensitive species such as Na and Ca ( Lagrange et al. 1998 ), which also seems to disfavour the preferential depletion scenario. The preferential production scenario, also proposed by Xie et al. (2013) , suggests that the overabundance is the result of gas being produced from materials rich in C and O. This includes photodesorption from C and O-rich icy grains ( Grigorieva et al. 2007 ) and collisional vaporisation of the dust in the disc ( Czechowski Mann 2007 ). Unlike preferential depletion, this scenario does not infer high N abundances and is thus more compatible with the near-solar NI abundance measured here. 5.2 Origin and dynamics of the NI gas The absorption components that we used to model the shape of the NI absorption lines may provide us with hints at both the origin and dynamics of the NI gas. The two absorbing circumstellar components, CS 0 and CS X , were not constrained to particular radial velocity values, but provided the best fit when their radial velocities matched that of the star. This indicates an absence of detectable amounts of NI falling in towards the star. The clearly asymmetric HI Ly- α line presented in Wilson et al. (2017) , on the other hand, indicates that HI gas is falling in towards the star. The HI gas is thought to be originating from the dissociation of water molecules from evaporating exocomets. The fact that we do not see similar asymmetric line profiles in NI is perhaps not surprising. In solar system comets, nitrogen-bearing molecules are only minor constituents ( Eberhardt et al. 1987 ), which means that if β Pic comets are similar, we would not expect the NI lines to be asymmetric. This does not exclude the possibility that NI might originate from exocomets, however, as we argue in Sect. 5.3 . 5.3 Role of exocomets The NI column density measurements are dominated by the narrow absorption feature, CS 0 , which has a line width of km s −1 . With a similar line width observed in FeI ( Vidal-Madjar et al. 2017 ), this absorption feature likely traces a stable band of circumstellar material that orbits β Pic. The origin of the broad absorption feature, CS X , with km s −1 , which absorbs seven times less strongly, but clearly governs much of the absorption line shape, is less clear. The feature might consist of a number of individual absorption components that are beyond the resolving powers of the COS instrument. The broad component could be related to NI absorption lines in the stellar atmosphere of β Pic. This would certainly explain that the bulk velocity matches that of the star. However, we would then expect the lines to be much broader and shallower
Paper: 2018MNRAS.478..197P (fulltext)
coupling, resonance, and weak inner planet coupling. We elaborate on these in the following subsections. Figure 2. RMS values of the eccentricities of the planets in a fiducial two-planet system under the influence of an eccentric perturber. The perturber has initial eccentricity e p = 0.1 and has an orbit co-planar with the inner system. The other parameters are the same as Fig. 1 . The solid coloured points represent the results of numerical integrations using secular equations, while the solid coloured curves are calculated using linearized theory; the dashed coloured lines show the forced eccentricity (equation 26 ). The dashed vertical line corresponds to ε 12 = 1. 2.3.1 Resonance When ω 1 ≃ ω 2 (note that this is approximately equivalent to the condition ε 12 ∼ 1), a potential resonance feature arises where large eccentricities can be excited in the inner planets, even for small e p . If we take ω 1 = ω 2 , equations ( 34 ) and ( 35 ) becomes We see that the eccentricity of planet 1 is boosted while the eccentricity of planet 2 is dampened near the resonance. The resonance feature is most pronounced when L 2 ≫ L 1 . To illustrate this, let L 1 = 0, then the forced eccentricity on the inner planet becomes If both the inner planets have zero initial eccentricity, then their eccentricity evolution is given by In this limiting case, according to the linear theory, at ε 12 ≃ 1 the eccentricity of the inner planet can become arbitrarily large, even for small initial values of e p . In reality, the linear theory breaks down as e 1 becomes too large, and higher order terms will keep e 1 to a modest value. An illustration of the resonance behaviour can be seen in Fig. 2 . Systems with larger ratios of m 2 / m 1 tend to exhibit pronounced resonance features, whereas for systems with more comparable masses, the feature is notably reduced. 2.3.2 Strongly and weakly coupled regime In the case where the mutual precession rates of the inner planets dominate over the influence of the perturber (i.e. ω 12 ≫ ω 1 p and ω 21 ≫ ω 2 p ), the general expressions ( 34 ) and ( 35 ) can be significantly simplified; we refer to this as the strongly coupled regime. In this regime, the two planets attain very similar forced eccentricities (), and as a result equations ( 34 ) and ( 35 ) are dominated by their first terms. Explicitly, in this regime the RMS eccentricities are approximately given by where α 12 = ( a 1 / a 2 ) and β 12 = β( a 1 / a 2 ) (see equation 8 ). In the other limiting case (i.e. weak coupling), when the precession rates of the inner planets driven by the perturber dominate the their mutual precession rates (i.e. ν 1 p ≫ ω 12 and ν 2 p ≫ ω 21 ), the terms of order ν 12 /ω 1 p and ν 21 /ω 2 p in equations ( 34 ) and ( 35 ) can be dropped, and the final inner planets’ RMS eccentricities are again given by . In this case, the planets precess independently of one another, and their eccentricities are given by equations ( 11 ) and ( 12 ), i.e. Note
Paper: 2017MNRAS.471.2790M (fulltext)
described by a modified lognormal distribution (equation A7), whose parameters are given in Table 3 . The density PDF converges after the jet breaks out (jet head ≳ 3 kpc) and decouples from the ISM, proceeding along the cleared path. Figure 7. The evolution of the density PDF as the jet evolves with time and breaks out of the dense central region after a height of z ∼ 3 kpc. The density PDF converges after the jet breakout as the jet decouples from the ISM. Table 3. Coefficients of the fit to the density PDF in Fig. 7 following equation (A7). a Mean density for the PDF. b Standard deviation of variable ρ, which is related to the standard deviation in s = ln ρ as in equation (A8). c Volume-weighted mean of density for n > 10 cm −3 . This gives a measure of the mean density of the highdensity filaments. The duration of the flood-channel phase of the jet evolution depends on the mean density and filling factor of the ISM, as previously discussed in Wagner et al. ( 2012 ). Clouds with denser cores are less ablated and more efficiently impede the progress of the jet. In Fig. 8 , we present the density for Sim B (Table 2), where a jet of power 10 45 erg s −1 passes through a medium with a lower mean density ( n w 0 = 150 cm −3 ) compared to that of Sim A. We note that the jet breaks out of the warm dense gas much faster, in comparison to Sim A. Figure 8. Density {log [ n (cm −3 )]}in the x – z plane for Sim B. 4.2 Evolution of the energy bubble The confined jet creates a nearly homogeneous high-pressured energy bubble that sweeps through the ISM (see Fig. 6 ). The expanding bubble engulfs the dense filaments, driving shocks into them. The shocks become radiative, forming a skin of heated gas (∼10 5 K) surrounding the warm filaments at ∼10 4 K. The energy bubble also shears the filaments sweeping up the ablated gas to high velocities ≳1000 km s −1 . The dense filaments are themselves accelerated by the ram pressure from the bubble. However, the efficiency of acceleration depends on the density of the gas and injected mechanical power. For lower density (Sim B), the dense filaments are accelerated to speeds of ∼100 km s −1 , similar to earlier reported results of Wagner et al. ( 2012 ). In simulations with higher ISM density (Sim A) or lower jet power (Sim C and Sim D), the outward pressure from the bubble is not sufficient to overcome the inward gravitational acceleration. In Figs 10 and 11 , we show the evolution of the mean pressure inside the energy bubble and the volume of the bubble. Initially, the bubble pressure decreases more rapidly ( p ∼ t −1 ) than an adiabatic spherical bubble (with p ∼ t −0.8 ). This is likely the result of mixing with the ISM and cooling losses. Consistent with the lower pressure, the bubble expands less rapidly (volume ∼ t 1.5 ) than an adiabatic bubble (volume ∼ t 1.8 , see Appendix B). At later stages after the breakout from the confining dense ISM, the bubble expands more rapidly. 4.3 Evolution of the phase space
Paper: 2019ApJ...887..226K (fulltext)
the analytical solution without collisional damping is in good agreement with the results of the simulations. The timescale of dynamical friction is estimated from Equations ( 14 ) and ( 15 ). If and ≫ for m p,1 ≫ m p,2 , the damping timescale due to dynamical friction has the relation as (Ohtsuki et al. 2002 )where M disk is the total mass of a planetesimal disk, is the width of the disk, and is the Hill radius for protoplanets. Note that P DF is estimated to be ∼10 for m p,1 ∼ M ⊕ and ≈ ∼ 0.1. This estimate implies a planetesimal disk with M disk ≪ M ⊕ leads to eccentricity damping for Earth-sized protoplanets in a long timescale ≫10 4 yr via dynamical friction. In addition, dynamical friction increases of planetesimals. The increase in makes τ df longer. On the other hand, the decreasing timescale for M disk due to the collisional cascade, τ cc , is estimated from Equation ( 5 ) to be (Kobayashi Tanaka 2010 )For 100 km sized planetesimals ( m c ≈ 10 22 g) with ∼ 0.1, τ cc ≲ τ df . It should be noted that dynamical friction increases during the damping, which shortens τ cc and elongates τ df as discussed above. Therefore, even if τ cc ∼ τ df initially, τ cc eventually becomes much shorter than τ df . Equations ( 17 ) and ( 18 ) show τ df / τ cc is independent of M disk and Δ a but depends on , , and m c . For = = 0.03, P DF ≈ 60 so that τ cc ≪ τ df for g, while τ df ∼ τ cc for g. Therefore, as shown in Figures 4 and 5 , e and i of protoplanets insignificantly decrease for m c = 10 16 g because of τ cc ≪ τ df , while those are moderately damped for τ cc ∼ τ df with m c = 10 19 g. Therefore, the condition with t cc ≫ τ df is required for the e damping of protoplanets. Planets formed via giant impacts have high eccentricities and inclinations. For Earth-sized planets formed in the giant impact stages, the mean eccentricity is ∼0.1 (Chambers 2001 ; Kokubo et al. 2006 ), which is much larger than the current eccentricities of Earth and Venus. The depletion timescale of a planetesimal disk with is shorter than the e damping timescale via dynamical friction (see Equations ( 17 ) and ( 18 )). Therefore, larger planetesimals are required for e damping of Earth-sized planets. We consider such large planetesimals as being produced from a giant impact. The mass ratios of the largest giant impact ejecta to parent protoplanets are ∼0.01 (Genda et al. 2015 ). Therefore, we consider g, resulting in τ cc ≫ τ df . We perform the simulation for the orbital evolution of a high-eccentricity planet with mass M ⊕ in a swarm of giant impact ejecta. Giant impact ejecta initially have similar orbits to parent planets. However, the orbital phases of giant impact ejecta, π and Ω, are eventually distributed uniformly due to perturbation from other planets. The timescale to achieve a uniform distribution for π and Ω, which is roughly given by the precession rates for π and Ω obtained from the secular perturbation theory (Murray Dermott 1999 ), is estimated as ∼10 5
Paper: 2017MNRAS.469.1259B (fulltext)
BCGs at 0.5 z 1.0 to derive L TIR and the associated dust-obscured SFR, dust temperature and mass, but neither attempt to mitigate confusion noise in these bands, nor match the observed far-infrared and 24 m emission in their SED template fitting (which are not simultaneously fit for significant number of their sources).] Furthermore, the results of the template fitting show that a customized combination of components provides a better description of the BCG stacked fluxes in most instances than any single template from a pre-existing library, particularly at z 1. It is clear that an elliptical template alone is not sufficient to match any of the SEDs, as the far-infrared emission of the BCGs is one to two orders of magnitude brighter than that of a typical red and dead elliptical galaxy template. As for the star-forming fit, based on the dust scaling relations empirically derived from observations of Herschel Reference Survey galaxies of a large range of spectral type, environment and stellar mass in the local Universe (Cortese etal. 2012 ) specifically the inverse proportionality between dust-to-stellar mass ratio and the stellar mass and specific star formation rate (sSFR) in local galaxies we expect the BCG SED to exhibit a significantly lower dust-to-stellar mass ratio in comparison to a less-massive non-BCG star-forming template, which is exactly what we find. At the lowest redshifts we consider, z avg 0.185, a typical star-forming SED matches the BCG SED longward of 10m, but the BCG stellar bump rises significantly above that of the star-forming template. While these dust scaling relations were derived for local field and cluster galaxies (including those in cluster cores), we do not yet know if they apply to galaxies at higher redshifts (though we might expect them to if we assume an understanding of the physics underlying the trends). Whereas massive elliptical galaxies in the local Universe exhibit relatively low dust-to-stellar mass ratios on account of a diminished level of star formation activity with respect to local star-forming late-type galaxies, we would expect an increasing dust-to-stellar mass ratio for elliptical galaxies of a fixed stellar mass with increasing sSFR towards higher redshift, as observed in our sample (reminding the reader that we select BCGs from clusters of the same mass at every redshift, therefore the BCGs are of the same mass at every redshift based on the linear scaling of BCG stellar mass and host-cluster mass, e.g. Lidman etal. 2012 ). In fact, as we move towards higher redshift, we observe the star-forming templates begin to match the full extent of the BCG SED better and better, as opposed to just the far-IR bump at lower redshifts. Therefore, we fit the stellar emission shortward of 10m in the faint SEDs with only the Polletta 13-Gyr Elliptical model, but add to this model either of the Polletta M82 or S18 templates to fit the stellar emission in the bright SEDs. In the mid- and far-infrared
Paper: 2018ApJ...869...26W (fulltext)
i.e., α 0 = −0.5 and α a = 0. In Figure 3 , we show T 21 , as defined in Equation ( 2 ), for various values of α 0 and α a , with other cosmological parameters fixed at values consistent with a Planck 2015 cosmology (Planck Collaboration et al. 2016 ). As shown, the ΛCDM model ( α 0 = α a = 0), denoted by the star, is in tension with the EDGES measurement at a 99% CL illustrated by the hatched region. However, interacting vacuum models can in principle release the tension, namely, T 21 can be pushed into the hatched region by a large range of the α 0 and α a parameters. Figure 3. Illustration of the intensity of the 21 cm signal relative to the CMB temperature, T 21 (K), for various values of α 0 and α a . The color bar indicates the values of α a . The black solid curve corresponds to the case in which α does not evolve with time ( α a = 0). The intersect between the black solid curve and the vertical gray dashed line, marked by a yellow star, denotes the ΛCDM model. The hatched region illustrates the observed 21 cm signal from EDGES at a 99% CL. We then perform a Monte Carlo Markov Chain global fit for the parameters in Equation ( 12 ) using a modified version of CosmoMC 8 8 Available at https://cosmologist.info/cosmomc/ . (Lewis Bridle 2002 ) with a combined data set including, 1. the angular power spectra of temperature and polarization measurements of CMB from the Planck mission (Planck Collaboration et al. 2016 ); 2. the Joint Light-curve Analysis sample of supernovae (SNe) measurements (Betoule et al. 2014 ); 3. the Baryonic Acoustic Oscillations (BAO) distance measurements from 6dFGS (Beutler et al. 2011 ), Sloan Digital Sky Survey (SDSS) DR7 Main Galaxy Sample (Ross et al. 2015 ), Ly α forest of BOSS DR11 quasars (Font-Ribera et al. 2014 ; Delubac et al. 2015 ), and BOSS DR12 with tomographic information (Wang et al. 2017 ; Zhao et al. 2017 ); joint BAO and Redshift Space Distortions (RSD) measurements from WiggleZ (Blake et al. 2012 ) and from eBOSS DR14 (Zhao et al. 2019 ); and RSD measurements from 6dFGS (Beutler et al. 2012 ), 2dFGRS (Percival et al. 2004 ), SDSS LRG (Samushia et al. 2012 ), and VIPERS (de la Torre et al. 2013 ); 4. and the local H 0 measurement using the Cepheids, i.e., H 0 = 73.24 ± 1.74 km s −1 Mpc −1 (Riess et al. 2016 ). The results are summarized in Table 1 and in Figures 4 and 5 . To quantify the constraint on α 0 and α a from the EDEGS measurement, we show the 68%, 95%, and 99% CL contours of α 0 and α a using EDEGS alone in the lower left part of Figure 4 . 9 9 As EDEGS alone cannot constrain all of the parameters in Equation ( 12 ) simultaneously, we only vary α 0 and α a in this case while the other parameters are fixed to the values derived from the Planck 2015 measurement. As illustrated, the EDGES measurement favors an interacting vacuum model over the ΛCDM model at more than 99% CL, and this would induce a quintessence-like effective dark energy as shown in the lower left of Figure 1 . We then
Paper: 2018ApJ...856..110Y (fulltext)
( 2010 ), 8 8 The quantity f end is set to avoid higher PN tidal effects and nonlinear hydrodynamical coupling. In Hinderer et al. ( 2010 ), the fitting formula was obtained by choosing 400 Hz ≥ f end ≥ 500 Hz. The end frequency adopted in Abbott et al. ( 2017c ) is the frequency of the innermost stable circular orbit, which is above 1 kHz. and an equal mass binary with M ∼ 2.7 M ⊙ , we obtain an estimate on the tidal deformability of a star: ΔΛ 1 (1.35 M ⊙ ) ∼ 2.9 × 10 3 . Assuming a low-spin prior, GW170817 places on upper bound Λ 1 (1.4 M ⊙ ) ≤ 800 at the 90% confidence level, and an upper bound of ≤1400 if the prior is relaxed to allow for high spin (Abbott et al. 2017c ). Such constraints on Λ 1 also limit the allowed radius of NS, given the family of EOS assumed in this paper. In Figure 3 , we compare this uncertainty in the tidal deformability to the amount by the tidal deformability that has to be changed in order for a BH–NS binary to have the same leading order tidal effects as a corresponding NS–NS binary with the same masses. Figure 3. Amount by which the dimensionless tidal deformability Λ has to be increased, relative to the quantity of the corresponding NS in an NS–NS binary with the same masses, for a BH–NS binary to have the same leading order tidal effects. This is shown as a function of the mass ratio. In all of the cases, one NS is assumed have a mass of M NS1 = 1.4 M ⊙ . The different curves correspond to different choices of EOS for the NS–NS binary, and hence the different values of the radius of 1.4 M ⊙ NS in the NS–NS, as shown in the legend. For comparison, we also show the Fupper bound constraints on Λ from GW170817 assuming either a high-spin prior or a low-spin prior on the source. The highest NS radius shown (dotted red curve) saturates the former constraint on Λ, while the middle value for the NS radius (dotted blue curve) saturates the latter constraint. With N identical detections, and under the same high-spin prior, such uncertainty scales as Δ N Λ 1 (1.4 M ⊙ ) ∼ 1.4 × 10 3 N −1/2 . In reality, the component masses and source distances are different for different events, and it is possible that the best event of previous N detections dominates the constraint of NS EOS. In the above discussion, we have not accounted for the effect of spin, which is particularly important for LMBHs formed through hierarchical mergers, as they are expected to have relatively high spin ( a ∼ 0.7; Fishbach et al. 2017 ; Gerosa Berti 2017 ). 9 9 The effective spin of a most rapidly spinning pulsar known so far is ∼0.4 (Hessels et al. 2006 ). Such a spin magnitude will generate ∼0.15 mismatch between a nonspinning BH–NS waveform and a generic pretcessing waveform if the mass ratio is around 2 (see Figure 8 of Harry et al. 2014 ). The relation between distinguishable mismatch and S/N is discussed in Baird et al. ( 2013 ). 5. Multi-messenger Detection An important question is whether multi-messenger signals can help us identify an
Paper: 2017A&A...599A..78P (fulltext)
L5 dwarfs includes 10 L4.5, 24 L5, and 2 L5.5 sources with seven detections of H in emission, and three classified as variable ( Schmidt et al. 2007 , 2015 ). In their sample, Pineda et al. (2016) included three L5 with H in emission in one of them. They inferred a fraction of emitters of 20% (15–35%; 68% confidence limit), consistent with the interval from the SDSS sample ( Schmidt et al. 2007 , 2015 ) although to date there are few sources. We also note that a large number of L dwarfs show variability in their level of chromospheric activity ( Liebert et al. 2003 ; Reiners Basri 2008 ; Schmidt et al. 2015 ; Pineda et al. 2016 ), the most striking case being the L5.5eT7 binary 2MASSI J1315309264951AB (L5.5; Hall 2002a ; Gizis 2002 ; Burgasser et al. 2011 , 2013 , 2015 ). 4.3. Radial velocity and space motion We measured the position of the H line in the 1D spectrum with the splot task under IRAF. We identified the line at 6564.11 Å with an error of 0.04 Å, while the nominal air wavelength from the NIST Atomic Spectra Database is 6562.819 Å. We should add in quadrature the error on our wavelength calibration (2.3 km s, see Sect. 3.2 ) and uncertainty due to the timespan in our observations (0.2 km s). We inferred a radial velocity of 38.0 2.9 km s (Table 2 ) after correcting for the rotation of the Earth, the motion of the Earth around the Earth-Moon barycentre, the orbit of the barycentre about the Sun, and motion of the Sun relative to the specified standard of rest. Our value is in close agreement with the mean values derived independently from the ground (39.1 0.2 km s; Detweiler et al. 1984 ) and from space with H ipparcos (39.48 0.30 km s; de Bruijne et al. 2001 ). We infer a space motion of = [42.92, 23.96, 0.72] km s for 2M041821 (Table 2 ), in close agreement with the space motion of the cluster centre: [41.1, 19.2, 1.4] km s ( van Leeuwen 2009 ). According to Vizier, 2M041821 should have a Gaia -band of 19.096 and should be included in the Initial Gaia Source List ( Smart Nicastro 2014 ), implying that precise coordinates and proper motions will be soon available since the second Gaia data release is expected at the end of 2017. However, based on the equation 0.244 SpT 12.6332 kindly provided by R. Smart, 2M041821 might have 20.4 mag, placing it at the faint end of the Gaia catalogue. Therefore, it is unclear at this stage whether its membership to the Hyades cluster will be settled by Gaia in the near future. Table 2 Results of our spectroscopic analysis. Notes. We quote the spectral type, spectroscopic distance, radial velocity, space motion, and ratio of the (/) luminosity for 2M041821. 4.4. Lithium absorption The lithium test was proposed in the early 1990s to distinguish between very low-mass stars and brown dwarfs ( Rebolo et al. 1992 ; Magazzu et al. 1993 ). The technique relies on the detection of lithium in absorption at 6707.8 Å for late-M and L dwarfs straddling the stellar-substellar boundary. If lithium is detected,