adsabs/WIESP2022-NER huggingface dataset:
Paper: 2018A&A...612A..30B (fulltext)
are often referred to as pebbles . Pebbles interact with the gas disc through gas drag and drift inwards ( Weidenschilling 1977 ; Brauer et al. 2008 ). Pebbles can become concentrated in pressure bumps and through the streaming instability (see Johansen et al. 2014 for a review), leading to planetesimal formation by gravitational collapse of the filaments. Planetesimals formed by the streaming instability have characteristic sizes of 100 km ( Johansen et al. 2015 ; Simon et al. 2015 ). In classical planet formation models, the cores of the giant planets form through mutual collisions between these planetesimals ( Pollack et al. 1996 ). However, to achieve a core mass high enough (≈10 M E ) to attract a gaseous envelope, a surface density of planetesimals of a few times the Minimum Mass Solar Nebula (MMSN) is needed. Additionally, the growth timescale increases steeply with orbital distance, making the formation of the ice giants in the solar system basically impossible to achieve with planetesimal accretion alone. Gravitational stirring of the planetesimals by a set of growing protoplanets decreases the growth rates even more ( Levison et al. 2010 ). In recent years, a new paradigm of solid accretion has emerged: pebble accretion ( Johansen Lacerda 2010 ; Ormel Klahr 2010 ; Lambrechts Johansen 2012 ; Morbidelli Nesvorny 2012 ). When a pebble enters the planetary Hill sphere, it is subject to gas drag, which robs the pebble of angular momentum, resulting in an inward drift of the pebble onto the planet. When the largest planetesimals have grown to a few hundred kilometers in size by accreting other planetesimals, rapid pebble accretion allows further growth to cores of ten Earth masses well within the lifetime of the protoplanetary disc ( Ida et al. 2016 ; Visser Ormel 2016 ; Johansen Lambrechts 2017 ). A growing planet opens a partial gap in the protoplanetary gas disc, which influences the motion of solids in the disc ( Paardekooper Mellema 2006 ; Rice et al. 2006 ). Pebble accretion stops when the gap carved by the planet generates a pressure maximum outside of its orbit, which stops the inward flux of pebbles ( Morbidelli Nesvorny 2012 ; Lambrechts et al. 2014 ). This is referred to as the pebble-isolation mass. As the influx of pebbles is stopped, the planet’s gas envelope loses its hydrostatic support, and the envelope can then contract to form a planet with an extensive gaseous atmosphere ( Lambrechts et al. 2014 ). In the solar system, pebble isolation is a potential mechanism to explain the dichotomy between the ice and gas giants, with ice giants never reaching the pebble-isolation mass and hence not able to undergo gas accretion within the lifetime of the protoplanetary disc ( Lambrechts et al. 2014 ; Venturini Helled 2017 ; Frehlik Murray-Clay 2017 ). Lambrechts et al. (2014) used hydrodynamical simulations to infer a pebble-isolation mass given by where H ∕ r is the discs aspect ratio. Lambrechts et al. (2014) found a weak
Paper: 2018ApJ...852..139V (acknowledgments)
O.M. and P.V.D. would like to thank the Fermilab Theoretical Physics Department for hospitality. O.M. and P.V.D. are supported by PROMETEO II/2014/050, by the Spanish Grants SEV-2014-0398 and FPA2014–57816-P of MINECO and by the European Union’s Horizon 2020 research and innovation program under the Marie Skłodowska-Curie grant agreements No. 690575 and 674896. This manuscript has been authored in part by Fermi Research Alliance, LLC under contract No. DE-AC02-07CH11359 with the U.S. Department of Energy, Office of Science, Office of High Energy Physics. The United States Government retains and the publisher, by accepting the article for publication, acknowledges that the United States Government retains a non-exclusive, paid-up, irrevocable, world-wide license to publish or reproduce the published form of this manuscript, or allow others to do so, for United States Government purposes. This research used resources of the Argonne Leadership Computing Facility, which is a DOE Office of Science User Facility supported under Contract DE-AC02-06CH11357. An award of computer time was provided by the Innovative and Novel Computational Impact on Theory and Experiment (INCITE) program. This research is also part of the Blue Waters sustained-petascale computing project, which is supported by the National Science Foundation (awards OCI-0725070 and ACI-1238993) and the state of Illinois. Blue Waters is a joint effort of the University of Illinois at Urbana-Champaign and its National Center for Supercomputing Applications. This work made extensive use of the NASA Astrophysics Data System and arXiv.org preprint server.
Paper: 2018ApJ...863..166V (fulltext)
17 yr, Feng et al. ( 2015 ) found strong evidence for a second planet with orbiting with a period of 15.1 ± 1.1 yr. Like GJ 628 and GJ 649, GJ 849 has nearly solar [Ti/Fe] and kinematics that skew the age posterior to younger ages. Based on this, we estimate the age of GJ 849 to be . 5.2.11. GJ 876 The planetary system around GJ 876 is a benchmark system for studying the formation and migration of planets in compact systems. GJ 876 hosts a total of four known planets, three of which are in a Laplace 1:2:4 mean-motion resonance (Rivera et al. 2010 ). The resonance is chaotic but expected to be stable on timescales of at least 1 Gyr (Rivera et al. 2010 ; Martí et al. 2013 ). The old age we infer for GJ 876 ( Gyr) suggests that such chaotic resonances can be stable for several Gyr, assuming the planets migrated into this configuration soon after their formation (Batygin et al. 2015 ). Even though the innermost planet, GJ 876 d, is not in the resonant chain with the other three planets, our old age for GJ 876 has some interesting implications for its past migration. Originally, Rivera et al. ( 2010 ) estimated that planet d orbited with an unusually high eccentricity ( e = 0.207 ± 0.055) for a planet that orbits at only 0.02 au with a period of about 2 days. Recently, Millholland et al. ( 2018 ) and Trifonov et al. ( 2018 ) reanalyzed the system by fitting dynamical N -body models to existing and new RV data. Both analyses revised the eccentricity of planet d down to values more consistent with a circular orbit. Trifonov et al. ( 2018 ) reported a best-fit eccentricity of , while Millholland et al. ( 2018 ) reported a best-fit eccentricity of e = 0.057 ± 0.039. Using the Trifonov et al. ( 2018 ) estimate for the eccentricity, we constrain the minimum Q p of GJ 876 d to be >5 × 10 5 . Such a high Q p is surprising for an ∼7 M ⊕ planet. Our minimum Q p for GJ 876 d is an order of magnitude higher than the Q p of Uranus and Neptune and three orders of magnitude greater than the Q p of terrestrial bodies. A few scenarios could explain the nonzero eccentricity of GJ 876 d. One explanation is that the planet has a peculiar structure that is very inefficient at dissipating tidal energy. Another explanation is that GJ 876 d originally orbited with a larger semimajor axis where it gravitationally interacted with the outer three planets and, at some point in its history, fell out of a chaotic resonance with the outer planets and migrated in. A third scenario is that there is another, unseen planet in the system that is gravitationally interacting with GJ 876 d. We note that the eccentricity of GJ 876 d has undergone downward revision recently. If additional observations in future studies lead to further downward revision and the conclusion that the orbit is essentially circular, it may not be necessary to invoke one of the above scenarios to explain the orbit of GJ 876 d. In that case, our minimum Q p will be overestimated. However, given its proximity
Paper: 2017ApJ...848..114P (fulltext)
and our purity of is calculated for real clusters with . However, our completeness decreases (80%) in this mass range (), as shown in Licitra et al. ( 2016a ). When fixing the slope at unity, we obtain () and a scatter of ( dex) for the complete (published) catalogs. In this case, the difference in a for the two samples is negligible, . The small shift in normalization (∼0.2 dex) compared to the diagonal is expected, because lensing mass estimates are generally higher than X-ray masses (Zhang et al. 2008 ; Rasia et al. 2012 ; Simet et al. 2017a ). When leaving the slope as a free parameter, we find and , with a scatter of ( and , with a scatter of ) for the complete (published) catalogs. The incompleteness when using the published catalogs appears to bias our fit slope, which becomes much shallower than the diagonal. In Figure 10 , we show the mass–luminosity and mass–temperature relations. We apply a logarithmic linear fit, in the form: where , for the , for the , , and . For the mass–luminosity relation, we find and , with a scatter ( and , with a scatter ) for the complete ( published ) catalogs. For the mass–temperature relation, we find and , with a scatter ( and , with a scatter ), for the complete ( published catalogs). The relations obtained with the published catalogs show again shallower slopes. Our results are consistent with the expected deviations from self-similarity (Böhringer et al. 2011 ). We summarize our results in Table 4 . Table 4 Results of the Fit of the Weak Lensing Mass vs. X-Ray Mass and Mass Proxy Relations: Note. “CC” refers to the complete catalogs and “PC” to the published catalogs (see text for the catalog definitions and values of the pivot mass, luminosity, and temperature used in the fit of the scaling relations). 6. Discussion 6.1. Comparison to Previously Derived Mass–Richness Relations In this section, we discuss our results in the context of similar current studies. As stated before and shown in Licitra et al. ( 2016a , 2016b ), our richness estimator λ is defined in a similar way as the richness from redMaPPer (Rykoff et al. 2014 ). The redMaPPer richness is defined as , where is the probability that each galaxy in the vicinity of the cluster is a red-sequence member and are weights that depend on luminosity and radius. In this calculation, only galaxies brighter than and within a scale radius are considered. The radius is richness-dependent and scales as . The RedGOLD richness is a simplified version of . We constrained the radial distribution of the red-sequence galaxies with an NFW profile and applied the same luminosity cut and radius scaling as in Rykoff et al. ( 2014 ), but did not apply a luminosity filter. Unlike the redMaPPer definition, our richness is not a sum of probabilities. Those choices were made to minimize the scatter in the mass–richness relation. For redshifts , the difference is only of 5–15%, while it increases to 40–60% at , where the redMaPPer richness is systematically higher
Paper: 2016ApJ...825...42V (fulltext)
deficit due to a gap (or hole) opened by the secondary black hole migrating within the circumbinary disk. Given the morphological evidence that Mrk 231 is indeed a recent merger (e.g., Hamilton Keel 1987 ; Hutchings Neff 1987 ; Surace et al. 1998 ; Veilleux et al. 2002 , 2006 ), and the dearth so far of black hole binaries in the cores of quasars (e.g., Eracleous et al. 2012 ; Popović 2012 ; Shen et al. 2013 ; Liu et al. 2014 ; Runnoe et al. 2015 ), contrary to theoretical expectations (Begelman et al. 1980 ; Yu 2002 ; Merritt Milosavljević 2005 ), the binary black hole explanation for the peculiar SED of Mrk 231 is potentially a very exciting new explanation of the Cycle 19 data, so it deserves to be examined in detail taking into account the new constraints from the Cycle 21 data. As pointed out by Yan et al. ( 2015 ), numerical simulations of black hole binaries (e.g., Miller Krolik 2013 ; Farris et al. 2014a , 2014b ; Roedig et al. 2014 ) show that the accretion onto the secondary black hole is expected to vary on a timescale of the order of the black hole binary orbital period, t orb ∼ 1.2 years, or less. However, as described here and in Paper I , the UV continuum emission in Mrk 231 has been surprisingly constant over the last several years. There is no evidence for any flux variations in the FUV continuum emission between 2011 (Paper I ) and 2014 (this paper), or ∼3 × t orb , and in the NUV continuum emission between 1992 (Smith et al. 1995 ) and 2014 (this paper), or ∼18 × t orb . This is not to say that the UV continuum of Mrk 231 does not vary at all. In fact, there is some evidence that the continuum level near 1300 Å in the 1978 archival International Ultraviolet Explorer ( IUE ) spectra of Mrk 231 (published in Hutchings Neff 1987 ) was ∼1.5× higher than the continuum level measured in our recent COS FUV data. But, the binary black hole model has difficulties explaining the stability of the UV emission in recent years. It would imply a (unrealistically) steady accretion stream of gas from the inner edge of the circumnuclear binary to the mini-disk of the secondary black hole over timescales of ∼3–18 t orb . A potentially fatal hurdle of the binary black hole model is the stability of the Ly α profile over a period of ∼3 year (∼3 × t orb ). In this model, the mini-disk of the secondary black hole is expected to dominate the line emission in the FUV, and the large blueshift of Ly α , C iii ], and C iv would therefore be due to the expected large line of sight velocity of the secondary black hole as it orbits the primary black hole. However, if this were the case, one would expect the centroid of the Ly α profile to shift in velocity from –3000 to at least +3000 km s −1 over half an orbital period. This is firmly ruled out by the new Cycle 21 data. Another difficulty with the binary black hole model is the fact that the absorption lines are only observed in the NUV and optical, but not in the FUV. This means that the BALR must be
Paper: 2017ApJ...847L...5P (fulltext)
all slices. This signature clearly indicates the existence of an antinode near the footpoints of the fan loop. 3.3. Estimation of Loop Length and Velocity of the Oscillations The footpoint of the fan loop that was away from the active region was distributed, and therefore not seen clearly in normal-intensity images of 171 or 193 Å. Thus, it was not straightforward to measure the length of the fan loops. Moreover, the shape and appearance of fan loops also changed with time (see animations a and b). To estimate the length, we chose the frames where the fan loops were best seen in normal-intensity images. We chose several points along the visible segment of the fan loops and interpolated a cubic spline between them. The length of the interpolated curve should be approximately equal to the projected length of the fan loops. The orange curves in Figure 2 are the fitted spline curves that represent the projected length of fan loops at the location of three artificial slices. We found the length of the fan loops at the locations of S1, S2, and S3 to be 62, 74, and 54 Mm, respectively (see Table 1 ). Note that the estimated length is the projected length in the plane of sky. Assuming the length of the fan loop is the same in 171 and 193 Å, we estimated the phase velocity of oscillations in 171 (193) Å to be 75 (85), 83 (101), and 65 (91) km s −1 at the locations of S1, S2, and S3, respectively. The phase velocity of oscillations, v , are comparable to the speed of sound in 171 and 193 Å. 3.4. Temperature and Density of the Fan Loop We estimated the temperature and density of the fan loop using the automated differential emission measure technique as developed by Aschwanden et al. ( 2013 ). The temperature of the fan loops was found to be ∼0.7 MK, which is much cooler than the hot loops as observed by hot SUMER lines (Wang et al. 2002 ) and in the AIA 94 Å channel (Wang et al. 2015 ). We also observed that the electron density decreased along the loop. Since the temperature of fan loops is low, they appear brighter in the 171 Å channel and diffuse in hotter channels like 193 Å. We should point out that a fan loop may consist of several finer strands and we have not considered that scenario here. 4. Discussion and Conclusions We observe intensity oscillations in a non-flaring fan loop system as seen in AIA 171 and 193 Å images. The intensity variations were out of phase close to two footpoints of fan loops, and the amplitude of the intensity oscillations varied along fan loops at the location of artificial slices. The amplitude of intensity oscillations first decreased and then increased while moving from one footpoint to another along the fan loop. It should be noted that it is difficult to identify the differences between standing and propagating waves without spectroscopic signatures. Recently, Yuan et al. ( 2015 ) have performed forward modeling of standing slow magnetoacoustic waves in flaring loops. They have reported that the variation of
Paper: 2016MNRAS.460.3170W (acknowledgments)
We would like to thank the anonymous referee for thoroughly reading the manuscript and providing useful feedback. We acknowledge funding from NSF ACI-1036211, NSF AST- 1517593, NSF AST-1009781, and the BlueWaters PAID program. The B lue T ides simulation was run on facilities on BlueWaters at the National Center for Supercomputing Applications. SMW acknowledges support from the UK Science and Technology Facilities Council (STFC) consolidated grant ST/L000652/1. ERS acknowledges support from the UK Science and Technology Facilities Council (STFC) consolidated grant ST/L000733/1. CL acknowledges support from a UK Science and Technology Facilities Council (STFC) PhD studentship.
Paper: 2018MNRAS.476.2465O (fulltext)
in magnetospheric size for the Earth, both with solar rotation and age. We calculate the magnetospheric size considering the Earth's magnetic field to be constant with time (blue solid line) and varying with time according to the model by Zuluaga et al. ( 2013 ; purple dotted line). This model is constrained by magnetic field measurements showing a 50 per cent field strength for a 1 Gyr system estimated by Tarduno et al. ( 2010 ). We can see that both models differ only slightly, with the most significant differences occurring around 1–2 Gyr, showing a spread of just under 1 R M . In particular, we find at early ages, both models predict magnetospheric sizes of ≈ 3 R E . This is understandable, considering the weak dependence on the planetary magnetic field of 1/6, shown in equation ( 23 ). We find a small magnetospheric standoff distance of 3.0 R E for a young Earth 100 Myr old, with the same planetary magnetic field as today ( B p = 0.3 G). These standoff distances would be similar to extreme modern day events, such as the 2003 Halloween storm (Rosenqvist et al. 2005 ), caused by coronal mass ejections (CMEs). However, these transient CME events last only on the scale of hours, but at 100 Myr these magnetosphere sizes would have been typical. Increased solar activity at younger ages (e.g. CMEs) would further compress the young magnetosphere, resulting in even smaller standoff distances. Younger stars are expected to be more active and would be more likely to produce large and more frequent transient events such as CMEs. Figure 4. We observe trends in magnetospheric size with respect to rotation (left-hand panel) and age (right-hand panel). Blue x's: calculated standoff distances with a constant planetary magnetic field of 0.3G. The solar-Earth scenario as it is currently, given as ⊕ . Purple dashed line: This depicts the standoff distances of the magnetosphere if the planetary magnetic field varied according to models described in Zuluaga et al. ( 2013 ). 5 DISCUSSION 5.1 Global properties Since direct observations of mass-loss rate and radio emission from stellar winds are difficult to obtain, we can use models to help understand the physical processes behind these winds. These models can provide information on the strength and location of the wind emission, therefore aiding in observing these winds directly. Recently, there has been research into the mass-loss rates of solar-analogue winds and how they would effect an orbiting planet, including a young Earth. Although some work focuses directly on the Sun–Earth interaction (Sterenborg et al. 2011 ), others focus more on the stellar evolution of these types of stars (Wood et al. 2002 , 2005 , 2014 ; Cranmer Saar 2011 ; Matt et al. 2012 ; Fichtinger et al. 2017 ) and their effects on exoplanets (Vidotto et al. 2012 , 2013 , 2015 ; See et al. 2014 ; Zuluaga, Mason Cuartas-Restrepo 2016 ; See et al. 2017 ). The main basis of our models is the temperature–rotation relation we presented in Fig.
Paper: 2021MNRAS.501.1755P (fulltext)
of the solver in regions with steep gradients, can be used to mitigate this problem at the expense of making the method more diffusive. There are many different possible limiters, each trading stability and diffusion in a different way (e.g. Toro 2009 ). High-order slope estimation would be particularly challenging in our galaxy formation simulations. We model stellar feedback by directly modifying gas properties in cells surrounding young star particles, often resulting in sharp discontinuities within the density, velocity, and energy fields. We have chosen the MinMod slope limiter (e.g. Toro 2009 ) which is the safest choice in terms of suppressing instabilities, but consequently also the most diffusive. We find that, if we adopt a less diffusive slope limiter, the solution becomes unphysical; with the current implementation of feedback, we are hence required to adopt the MinMod slope limiter in order to maintain stability. In contrast, galaxy simulations accounting for strong stellar feedback via source terms in the Riemann solver, e.g. using a non-thermal pressure (Agertz et al. 2013 ; Kretschmer, Agertz Teyssier 2020 ), have successfully been carried out using slope limiters such as MonCen (Agertz Kravtsov 2015 ). We leave a more systematic investigation of this topic for future work, and instead focus here on a completely new solution to the diffusion problem. If neither the solver nor the cell size can easily be changed, the final alternative is to ensure the rate of flow relative to the grid is small. However, haloes with a naturally small net velocity are exceptionally rare. For example, in the region used by the EDGE project, of all haloes move at velocities greater than at the time of early star formation, z = 16. As we will see, this is unavoidable in cosmological volumes, and already enough to cause substantive numerical errors. Further, adding a spatially uniform velocity offset to the initial conditions cannot usefully reduce even a single halo’s net speed: the cosmological expansion ‘redshifts’ the momentum of particles. Any uniform velocity offset that is not associated with a potential gradient therefore decays proportionally to 1 + z ; in contrast, the physical velocities grow (during matter domination in the linear regime) as (1 + z ) −1 . Therefore, uniform velocity offsets rapidly become insignificant compared to the growing physical velocities. Our solution is to use the genetic modification approach (Roth, Pontzen Peiris 2016 ; Rey Pontzen 2018 ; Stopyra et al. 2020 ) to post hoc modify the initial conditions such that the velocity of a chosen halo is minimized, while its environment, accretion history and other properties are minimally affected. In fact, we have already adopted this approach as a matter of course for published EDGE results (Rey et al. 2019 , 2020 ). In this paper, we examine the effects of disabling the modification in one of our galaxies in order to compare the AMR solution with and without rapid