adsabs/FOCAL huggingface dataset:
(i) Our capability of reproducing the lensed images down to the noise level without fully correctly modelling the cored central mass density distribution of the lenses indicates some form of the source-position transformation (SPT; Schneider & Sluse 2014), in line with the previous findings by Unruh et al. (2017). As a consequence, our reconstructions lead to a systematic fractional error on the Hubble constant of $25_{-19}^{+37}$ per cent (in comparison to a statical error of $12_{-3}^{+6}$ per cent when the shape of the lensing potential is perfectly known). This result is in agreement with the latest analysis of Blum, Castorina & Simonović (2020) that shows that cored (dark matter) mass density distributions give rise to approximate MSDs, and an error on the inferred Hubble constant. The latest cosmographic analyses (see e.g. Wong et al. 2019) have attempted to break these degeneracies by including the information contained in the kinematic properties of the lens galaxies and the positions of the lensed quasar images. However, the validity of this approach has been recently debated by Kochanek (2020), who has demonstrated that departures from single power-law mass distributions are responsible for a fractional error on the Hubble constant of 30 per cent. While the cores in the simulations analysed in this paper are artefacts related to limited resolution, cored mass density distribution in real galaxies may be developed by the effect of baryonic processes (see e.g. Chan et al. 2015) or changes in the dark matter properties (Schive, Chiueh & Broadhurst 2014; Spergel & Steinhardt 2000). Moreover, similar additional complexities exist in real galaxies are related, for example, to the presence of faint discs (Hsueh et al. 2018), bars, or other (baryonic) structures (Gilman et al. 2018; Xu et al. 2013). More generally, there exist many plausible deviations from a smooth power-law distribution, such as broken power laws (see e.g. Du et al. 2020) or multiple component models (see e.g. Nightingale, Dye & Massey 2018), which can produce comparable degeneracies. Together with the findings of Blum et al. (2020), our results have important implications for the analysis of time delays and a potential solution to the H0 tension (Wong et al. 2019).
Citation: Wong et al. 2019
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In Fig. 1, the state of the corona is shown both before the introduction of Alfvénic perturbations (panels a–c), as well as after the solar wind has reached a quasi-steady state upon the injection of Alfvén waves of the lowest considered pump frequency of 0.3 mHz (panels d–f). The solar wind density, sound speed, and radial speed are shown at these two evolutionary instances for the three flux-tube geometries considered, with expansions of fmax = 3, 5, and 10. The response of a polytropic solar wind (γ  5/3) for different magnetic flux tube expansions in the absence of a pump wave has been described in previous works (Kopp & Holzer 1976; Pinto et al. 2016). Close to the Sun (r  ∼2 RS), the corona is seen to exhibit similar plasma conditions irrespectively of the chosen flux-tube geometry. The heliocentric distance to which these similar plasma conditions persist (r  ∼2 RS in our study) is partly due to the choice of parameters in Eq. (3). At greater heliocentric distances, the mass density is higher for smaller values of fmax, while the opposite is observed for the sound speed and radial velocity. Upon injecting a 0.3 mHz Alfvén wave, we see that the solar wind exhibits no visually discernible perturbations, indicating that this medium-frequency (0.3 mHz) Alfvén wave simply acts as a direct source of momentum (via the wave pressure gradient) and energy to the solar wind. As compared to the steady state, the sound speed and density are lower and radial speed higher for all considered fmax. This behaviour of the solar wind in response to hour-scale Alfvén waves is similarly found in other numerical simulations (Shoda et al. 2018a) and is indicative of the low dissipation of Alfvén waves at these frequencies either by reflections in the medium or by PDI. Such hour-scale Alfvén waves are also seen to dominate observations (Banerjee et al. 2009) as these waves do not dissipate and propagate through the medium as in Fig. 1. The acceleration of the solar wind is discussed in Sect. 4.2.
Citation: Pinto et al. 2016
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where WCRSF is the width of the cyclotron line, Fν is the emergent flux including cyclotron scattering, and F0 the input continuum flux without a CRSF. In the Monte Carlo simulations performed by Schwarm et al. (2017b), only 1–10% of the initial photons undergo resonant scattering, that is, r ≳ 90%. Consequently, we would expect τCRSF ≲ 0.1 for the observed optical depths9. From our best-fit model of the spectra of Epoch II, the calculated residual fluxes after Eq. (10) are ≥78% for all three detected cyclotron lines (see Table 2). These values together with the fitted optical depths of τCRSF ≤ 0.24 (see Table 1) are in excellent agreement with the expectations from the simulations by Schwarm et al. (2017b). Although spawned photons from higher harmonics, which originate from multi-photon scattering and from the radiative de-excitation of electrons excited into higher levels, can affect the shape of the fundamental line (Isenberg et al. 1998), they can only decrease its depth. We note that residual fluxes down to r ≳ 36% have been observed for GX 304−1 (Rothschild et al. 2017), which can be obtained by, for example, a higher optical depth at the line-forming region. However, the CRSFs are not expected to absorb a significant fraction of the total broad-band continuum flux, which is the case for the model by Iyer et al. (2015). We conclude that the CRSF parameters that these latter authors found are due to a degeneracy between the continuum modeling and the CRSF modeling when unphysical values for the CRSF are allowed. In summary, the choice of a phenomenological model for the X-ray spectrum of an accreting neutron star may cause strong discrepancies relative to the theoretically expected values for the CRSF parameters. In particular, an erroneous shape or an overestimation of the X-ray flux is sometimes corrected by the introduction of further strong absorption components, which are probably not real. We stress that we do not claim that our choice of the phenomenological continuum model for 4U 0115+634 describes its true X-ray spectrum best. However, in contrast to other continua the model applied here yields CRSF parameters which are consistent with theoretical expectations. These conclusions desperately call out for a self-consistent model for both the spectral continuum and the cyclotron resonant scattering features.
Citation: Rothschild et al. 2017
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Despite 20 years of deep submm surveys since Smail et al. (1997), our knowledge of the upper half of the redshift distribution of SMGs remains incomplete. Early attempts to determine redshifts were conducted toward SMGs with radio counterparts because low-resolution (sub)mm images obtained with single dishes require high-resolution radio continuum maps from radio interferometers such as the Karl G. Jansky Very Large Array (VLA) in order to pinpoint source positions (Ivison et al. 1998, 2000, 2002, 2005, 2007; Smail et al. 1999; Borys et al. 2004; Pope et al. 2006; Aretxaga et al. 2011; Biggs et al. 2011; Yun et al. 2012; Umehata et al. 2014). Intensive studies of radio-bright SMGs were able to yield spectroscopic redshifts for those out to z ∼ 3 (e.g., Chapman et al. 2003, 2005). However, at that time, radio sensitivities could not detect SMGs beyond z ∼ 3, and as many as half of the SMGs lacked reliable radio counterparts (see e.g., Ivison et al. 2007; Biggs et al. 2011, cf. Lindner et al. 2011). Later attempts to determine SMG positions and redshifts using near- and mid-IR imaging could not fully overcome the bias toward lower redshifts, since the K corrections there are no more favorable than those in the radio regime, such that high-redshift sources are much fainter (e.g., Wardlow et al. 2011; Yun et al. 2012). Millimeter (mm) spectroscopic surveys toward gravitationally lensed dusty star-forming galaxies, taking advantage of their apparent ultra brightness, revealed a redshift distribution stretching out to z ∼ 5.8 (e.g., Vieira et al. 2013; Weiß et al. 2013; Strandet et al. 2016). These surveys suggested a larger fraction of SMGs at z ≳ 3 than previous studies of unlensed SMGs, perhaps partly because they were selected at 1.3 mm rather than the traditional 0.8–1.1 mm, but also because the requirement for high magnification favors galaxies with a long line of sight. We need to reveal the intrinsic redshift distributions of unlensed SMGs in large contiguous maps to determine their abundance in the early Universe and to study the evolution of the most massive galaxies via abundance matching with other populations, and with cosmological predictions (e.g., Hayward et al. 2013; Cowley et al. 2015).
Citation: Pope et al. 2006
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We can also compare to the models by Dessart, Audit & Hillier (2015), who synthesize light curves and spectra for superluminous SNe IIn with a variety of physical CSM parameters: in their fig. 20, they show the simulated bolometric light curves for their models out to +450 d. Past 200 d, most of the light curves experience a sudden and significant drop in luminosity that we did not observe for SN 2009ip, except for three (xe3m6r, xm3 and xm6). These models are distinguished from the others by having progenitor stars that experienced a more rapid mass-loss rate during the late stages of stellar evolution, and in one case, a CSM that extends to a radius 50 per cent farther than the others. The total bolometric luminosities for these models are ≳3 orders of magnitude higher at late times than SN 2009ip (log Lbol ≈ 40; Fraser et al. 2015) owing to a more energetic underlying explosion driving the interaction. The synthesized early-time spectra for these three models are also qualitatively similar to those of SN 2009ip (see fig. 22 of Dessart et al. 2015). Late-time synthesized spectra are not shown for these three models, but are given for their model X, in a comparison to luminous SN IIn 2010jl (see also Section 3.4). The synthetic spectra exhibit significantly more asymmetry and have a strong, blueshifted component of the H α line, which is even more exaggerated than that seen for SN 2010jl at ∼200 d. Dessart et al. (2015) explain this blueshifted feature as arising from an optically thick, cool dense shell between the reverse and forward shocks in the CSM, but for SN 2009ip we would not expect to see this signature because we found that the CSM exterior to the shock front is not yet optically thin (i.e. Section 3.2.1). Although not every aspect of the models by Dessart et al. (2015) represent the physical scenario of SN 2009ip, where appropriate comparisons can be made they support a moderately energetic explosion into a large and extended mass of CSM.
Citation: Fraser et al. 2015
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The most natural explanation is that the Compton-thick nature of some low-luminosity AGNs results in an underestimation of the true LX(2−10 keV) of these sources. These LX(2−10 keV) estimates come from studies using the spectra of LINERs at energies below 10 keV. However, Compton-thick sources show the bulk of the AGN power at energies above 10 keV. The intrinsic luminosity could be 10−70 times higher than the estimated using only energies below 10 keV in the Compton-thick scenario (Maiolino et al. 1998). González-Martín et al. (2009b) classified around 50% of their LINER sample as Compton-thick candidates. Figure 8 shows the AGN-dominated objects in our study attending to their Compton-thin (left) and Compton-thick (middle) classification. Most of the Compton-thin sources are close to the previously reported correlation for AGN. Compton-thick S2s are nicely placed along the AGN correlation found by Asmus et al. (2011). This is expected because the X-ray luminosities included for S2s are all corrected for their Compton-thick nature (most of them included in Goulding et al. 2012), either using X-ray measurements above 10 keV or assuming a factor between the observed and intrinsic X-ray luminosity for other Compton-thick AGNs (see Panessa et al. 2006, for details in this conversion factor). Compton-thick LINER candidates tend to be shifted towards X-ray luminosities lower than predicted for the AGN correlation. However, most of them are not consistent with the SB correlation either. The double arrows of Fig. 8 (middle panel) show their expected locus if the X-ray luminosity were ~10−70 times higher. Most of the Compton-thick LINERs can be placed in the AGN correlation if the correction is applied. The linear fit to the entire sample, once the intrinsic X-ray luminosity is corrected, results in LX(2−10 keV)(intrinsic) = 40 × LX(2−10 keV)(observed) for Compton-thick LINERs (marked as white stars in Fig. 8, right panel), which is very close to the linear relation found for AGN. Nonetheless, four LINERs and one S2 still remain very close to the SB correlation.
Citation: González-Martín et al. (2009b)
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A common property of observed transverse coronal loop oscillations is that they are damped quickly, usually within a few oscillation periods. It is now generally accepted that the main damping mechanism is resonant absorption (Sakurai et al. 1991; Goossens et al. 1992, 2002; Ruderman & Roberts 2002), transferring energy from the global kink mode to local azimuthal Alfvén modes in the boundary layer of the loop structure, where the two frequencies match (for a review of theoretical results, see Goossens et al. 2011). Less frequently, nearly undamped or even growing transverse oscillations are observed (Wang et al. 2012). Decay-less low amplitude kink oscillations, which are present in loops before and after a high-amplitude flare triggered damped kink oscillations can be explained in terms of a damped linear oscillator excited by a continuous low amplitude harmonic driver (Nisticò et al. 2013; Anfinogentov et al. 2013). Some examples of observed undamped high amplitude coronal loop oscillations can be found in Aschwanden et al. (2002) and Aschwanden & Schrijver (2011). In the latter paper, it was concluded that this could only happen if the thickness of the boundary layer (a radially inhomogeneous outer layer of the coronal loops) is much smaller than the loop radius, in order to minimize the damping due to resonant absorption. It was also evident from the observations that the loop was cooling during the particular undamped oscillation event. It seemed a rather natural explanation that the amplification due to cooling may counterbalance the damping due to resonant absorption, explaining the undamped oscillations, as it was first suggested in Ruderman (2011b). The most elaborate analytical study regarding time-dependent kink oscillations (Ruderman 2011a) considered the simultaneous effects of both damping due to resonant absorption and amplification due to plasma cooling. The conclusion was that, for typical boundary layer thicknesses, the amplification due to plasma cooling can account for the observed undamped oscillations only if the cooling happens quickly, on a timescale close to the oscillation period. However, in the paper, the effects of cooling and resonant damping have been studied under the assumption that both the characteristic cooling time and damping time are much longer than the characteristic oscillation period, when using the WKB method. Thus, it is questionable whether the derived equations remain valid for rapid cooling, i.e. for cooling times near to the oscillation periods. On the other hand, neglected non-linear behavior may change the outcome considerably, for example the Kelvin-Helmholtz instability at the tube boundary (Heyvaerts & Priest 1983; Ofman et al. 1994; Terradas et al. 2008), and the ponderomotive forces for high amplitudes (Terradas & Ofman 2004). The aim of this study is to further investigate the effects of radiative plasma cooling on the fundamental standing kink oscillation by means of numerical analysis.
Citation: Aschwanden et al. (2002)
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The most unique and characteristic observational feature of blazars’ highly variable broadband emission is the broad bimodal SED extending from the lowest accessible EM band, i.e., the radio, to the highest accessible, i.e., GeV-TeV γ-rays. The broadband SED of all blazars can be categorized into three different spectral subclasses: low-energy-peaked (LBL/LSP), intermediate-energy-peaked (IBL/ISP), and high-energy-peaked (HBL/HSP; Fossati et al. 1998; Abdo et al. 2010), based on the location of the low-energy hump. A remarkable property of each spectral subclass is the stability of the location of the two peaks despite huge variations in flux and often spectral shape. Only in a few rare instances has an appreciable shift in the location of the peaks been observed, e.g., the 1997 outburst of Mrk 501 (Pian et al. 1998; Ahnen et al. 2018) and the activity of OJ 287 from the end of 2015 to the middle of 2017 (Kushwaha et al. 2018a, 2018b). Even these two cases are remarkably different. In the case of Mrk 501, the locations of both the peaks shifted to higher energies. On the contrary, in OJ 287, a shift in the location of only the high-energy peak was observed during the 2015–2016 activity (Kushwaha et al. 2018a, 2019), while in 2016–2017 a new broadband emission component overwhelmed the overall emission, appearing as an overall shift in both the peaks as revealed in the detailed study by Kushwaha et al. (2018b). With the SED being the prime observable for exploration of the yet-debated high-energy emission mechanisms, such changes offer invaluable insights about the emission processes. For example, in Mrk 501 the shift in both peaks strongly implies the same particle distribution for the overall emission, while for OJ 287 the shift of only the high-energy peak can be reproduced by either inverse Compton scattering of the broad-line region photon field (Kushwaha et al. 2018a) or emission of hadronic origin (Oikonomou et al. 2019; Rodríguez-Ramírez et al. 2020).
Citation: Kushwaha et al. 2018a
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We present the photometry of the companion in Fig. 4 in a color-magnitude diagram. The corresponding numerical values are reported in Table 2. YSES 2b is consistent with a late L to early T spectral type when comparing it to colors of field brown dwarfs from the NIRSPEC Brown Dwarf Spectroscopic Survey (McLean et al. 2003, 2007), the IRTF spectral library (Rayner et al. 2009; Cushing et al. 2005), the L and T dwarf data archive (Knapp et al. 2004; Golimowski et al. 2004; Chiu et al. 2006), and the SpeX Prism Libraries (Burgasser et al. 2004, 2008, 2010; Gelino & Burgasser 2010; Burgasser 2007; Siegler et al. 2007; Reid et al. 2006; Kirkpatrick et al. 2006, 2010; Cruz et al. 2004; Burgasser & McElwain 2006; McElwain & Burgasser 2006; Sheppard & Cushing 2009; Looper et al. 2007, 2010; Muench et al. 2007; Dhital et al. 2011). Object distances were derived from Gaia EDR3 (Gaia Collaboration 2021), the Brown Dwarf Kinematics Project (Faherty et al. 2009), and the Pan-STARRS1 3π Survey (Best et al. 2018). In color-magnitude space, YSES 2b is very close to the innermost three planets of the HR 8799 multi-planetary system (Marois et al. 2008, 2010). These three planets are classified as mid to late L type dwarfs (e.g., Greenbaum et al. 2018), which agrees well with the sequence evolution of the adjacent field brown dwarfs from L to T spectral types3. A similar spectral type in this domain, therefore, seems very likely for YSES 2b, requiring confirmation by measurements at higher spectral resolution. Whereas the masses of the spectrally similar trio of HR 8799 c, d, and e are in the range 7–12 MJup (Wang et al. 2018; Marois et al. 2008, 2010), it is likely that YSES 2b has an even lower mass as the system age of (13.9 ± 2.3) Myr is significantly younger than the age of HR 8799, which is claimed to be member of the Columba association with an age of 30–50 Myr (Zuckerman et al. 2011; Bell et al. 2015). This is supported by the AMES-COND and AMES-dusty models (Allard et al. 2001; Chabrier et al. 2000) that we present in Fig. 4 for a system age of 13.9 Myr. An individual evaluation of these isochrones yielded masses from 5.3 MJup to 8.0 MJup as presented in Table 2. The uncertainties originate from the errors in the system age and planet magnitude that were propagated by a bootstrapping approach with 1000 randomly drawn samples from Gaussian distributions around both parameters. When combining the posterior distributions for the different models and filters we derived a final mass estimate of 

$6.3^{1.6}_{-0.9}\,M_{\mathrm{Jup}}$6.3−0.9+1.6 MJup
 as the 68% confidence interval around the median of the sample. This estimate is based on broadband photometric measurements alone; further spectral coverage of the planetary SED will be important to constrain its effective temperature, luminosity, surface gravity, and mass.
Citation: Marois et al. 2008
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Idealized simulations have the advantage of carefully controlled conditions but the disadvantages that the turbulence is driven in an artificially prescribed manner to maintain a fixed overall turbulent amplitude and the processes leading to cloud formation and destruction are not followed. In reality, GMCs form due to a combination of large-scale ISM flows (including turbulence, shear, and epicyclic motion) and gravity (both stellar gravity and self-gravity) that lead to collection of material from a large volume, as mediated by thermal and magnetic pressure, and a change from the atomic to the molecular phase as the gas cools (e.g., McKee & Ostriker 2007; Dobbs et al. 2014; Chevance et al. 2020). Turbulence on scales less than the scale height of the warm–cold ISM likely originates primarily due to the feedback from young stars (Elmegreen & Scalo 2004; Mac Low & Klessen 2004; McKee & Ostriker 2007),3

3
Gravitational instabilities in the combined gas–stellar system (e.g., Jog & Solomon 1984; Romeo 1992; Rafikov 2001; Kim & Ostriker 2007) can drive horizontal motions at very large scales, as seen in numerical simulations (e.g., Kim & Ostriker 2007; Shetty & Ostriker 2008; Agertz et al. 2009; Dobbs et al. 2011; Hopkins et al. 2012; Agertz & Kravtsov 2015, and citations within), but these motions generally do not reach supersonic amplitudes unless they are associated with gravitational collapse. In addition, turbulence at scales less than the disk scale height can be driven by spiral shocks and the magnetorotational instability, but numerical simulations show that the corresponding amplitudes are relatively modest in cold gas (e.g., Wada & Koda 2004; Piontek & Ostriker 2005, 2007; Kim et al. 2006, 2010; Dobbs & Bonnell 2007; Bonnell et al. 2013, and citations within).
 whether inherited from a GMC’s formation stage or produced internally. Considering that GMCs live for at most a few turbulent crossing or freefall times (Kawamura et al. 2009; Kruijssen et al. 2019), it is not clear that internal GMC conditions can control star formation in a way that is entirely divorced from their formation and destruction processes.
Citation: Dobbs et al. 2014
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