adsabs/WIESP2022-NER huggingface dataset:
Paper: 2017A&A...601A..34L (fulltext)
. 4.6. Perspectives with Gaia At the considered range of distances, Gaia will deliver exquisite parallax accuracy (%), especially because the binary effect can be adequately subtracted thanks to the knowledge of the astrometric orbital parameters presented in this study. Combined with the size of the apparent orbits and the periods, we expect the total masses to be measured with an uncertainty of 5%, following Eq. ( 1 ). For most systems, the error budget will be equally shared between the interferometric resolved astrometry and the Gaia distance. These measurements of the total masses will be completely independent from radial velocities. Gaia will also constrain the orbital photocentre displacement . This quantity can be extracted even if the orbital period is poorly sampled by Gaia because of the knowledge of all orbital parameters. It permits calculating the mass ratio : where is the (known) size of the apparent orbit and the flux ratio in the Gaia band. Interestingly, is well approximated by the (known) flux ratio in the band because hot stars have their optical and near-infrared spectra in the Rayleigh-Jeans regime. In order to test feasibility, we computed the expected photocentre displacement from the orbits derived in this study. This was done by inverting Eq. ( 5 ), guessing the expected mass ratio. We used two approaches, both relying on the calibrations from Martins et al. (2005) . We first converted the observed flux ratio into mass ratio. For a pair of main-sequence stars, the calibration is well approximated by . The exponent is even lowered for a massive, evolved primary with a less-massive main sequence secondary. In a second approach, we simply use the expected mass ratio from the individual spectral types when available (see Table 1 ). The computed are reported in Fig. 2 . The predicted displacements range from 0.1 mas to 1 mas. This is a factor ten to a hundred larger than the Gaia performances (assuming the orbital photocentre shift is measured with a similar accuracy than the parallaxes). Accounting for 5% uncertainty on the total mass (shared between the interferometric resolved astrometry and the Gaia distance, see Eq. ( 1 )), and for 1 to 10% uncertainty on the mass-ratio (shared between the interferometric flux ratio and the photocentre shift amplitude from Gaia , see Eq. ( 5 )), we expect to measure the individual masses with an uncertainty of %, independently of radial velocities. Fig. 2 Predicted photocentre displacement versus the flux ratio for the ten binaries of the sample. The mass ratios were estimated either from the flux ratio (assuming main sequence stars, filled symbols), or derived from the individual spectral types when available (open symbols). The dashed red line is for the expected Gaia performances. 5. Conclusion In this paper we report observations of spatially resolved astrometry on ten massive O-type binaries, using OLBI. This dataset is used to derive the elements of the apparent orbits. Combined
Paper: 2016ApJ...829...97H (fulltext)
the ICME structure and how the structure controls the geomagnetic activity. These results are summarized and discussed below, which illustrate the importance of multi-spacecraft remote-sensing and in situ observations for understanding the physical processes of CME propagation and space weather forecasting. 1. This study compares STEREO stereoscopic wide-angle imaging observations with multi-point in situ measurements at Mercury and Earth, which places a strong constraint on CME Sun-to-Earth propagation. The CME kinematics determined from the triangulation technique predicts well the shock arrival time at MESSENGER with an error of about 1 hr in this study. A reasonable accuracy is also obtained when we compare the predicted arrival time and speed with the in situ measurements near the Earth. From the Sun to Earth, the CME undergoes an impulsive acceleration, a rapid deceleration, and then a gradual deceleration out to 1 au, which agrees with the conclusions of Liu et al. ( 2013 , 2016 ). We find that the rapid deceleration ceases at ∼50 before the CME reached MESSENGER (∼0.47 au), which is different from the coronagraph findings in Gopalswamy et al. ( 2001a ). Combining this case with the three events in Liu et al. ( 2013 ), we suggest that fast CMEs are likely to terminate deceleration before reaching 100 from the Sun, which should be noticed in kinematics models of fast CMEs. This work once again proves the reliability of the stereoscopic triangulation technique in determining the kinematics of earthward CMEs in the inner heliosphere. 2. Our comparison between wide-angle imaging observations and the interplanetary type II radio burst indicates important clues on the source region of the type II burst. The study reveals that the consistency between the shock distance derived from the type II radio burst and the CME kinematics determined by the triangulation method requires an unusually high solar wind electron density. The type II radio burst was probably produced from a high-density region. This discrepancy and the slow increasing distance of the CME-driven shock derived from the type II radio burst can be explained by the shock interacting with nearby streamers (e.g., Reiner et al. 2003 ; Cho et al. 2007 ; Feng et al. 2012 ). Another possibility is that the type II radio burst was generated by the shock flank region with lower heights (Claßen Aurass 2002 ; Knock et al. 2003 ). Because only one spacecraft detected the type II radio burst, it is not possible to determine the position of the source region using the radio triangulation method of Martínez-Oliveros et al. ( 2015 ). This result implies uncertainties in the determination of CME kinematics using type II radio bursts alone. 3. We reconstruct the ICME structure near the Earth in order to connect it with the solar source observations and to understand how the structure contributes to the geomagnetic storm. The flux-rope inclination angle and chirality at 1 au are consistent with those
Paper: 2017MNRAS.470.3332I (fulltext)
43 , otherwise the matter is swallowed by the BHs before the merger. The bounding energy of this matter is only a fraction of the rest mass energy of the matter, This ratio is much smaller than the wind efficiency w 0.1 of a super-Eddington accretion disc, so that such matter is easily blown away by the disc wind. As long as a possible dead disc is ionized by the ISM accretion (that occurs unless we consider low n and high v ), we have encountered an unlikely set-up. Note that a fraction of the matter M m should accrete on to the BHs before the merger, otherwise a fine-tuning is needed because the time t b is much larger than the event duration t acc . The accretion is super-Eddington, even if only a fraction of the matter accretes, and should be accompanied by a strong disc wind as suggested by numerical simulations (Ohsuga etal. 2005 ; Jiang, Stone Davis 2014 ; Sadowski etal. 2014 ). Therefore, it is difficult to keep the matter near the BH before the merger and the BH mergers unlikely accompany observable prompt electromagnetic counterparts. 7 SUMMARY AND DISCUSSIONS We suggest possible connections between the BH mergers observed by GWs and the high-energy sources of TeVPeV particles in our Galaxy. The GW observations give a lower limit on the number of merged and hence highly spinning BHs as , and the spinning BHs produce relativistic jets by accreting matter and magnetic fields from the ISM. We calculate the luminosity function, the total power and the maximum acceleration energy of the BH jets, and find that the BH jets are eligible for PeVatrons, sources of CR positrons and TeV unIDs. The BH jets form extended nebulae-like PWNe. If they are observed as TeV unIDs, additional 300 nebulae will be discovered by CTA. We quantify the uncertainties of the estimate for the total power of the BH jets within a factor of 10 3 , which is much better than before the GW detections, by considering the initial BH spin, the velocity distribution depending on the formation scenario, the accretion profile changed by the wind and the feedback by the outflow (Table 2 ). The uncertainties will be reduced by the GW observations, in particular, of the BH spins. It is also important to clarify the feedback by the wind from the sub-Eddington accretion disc on the BondiHoyle accretion. Our considerations on the BH accretion and jet imply that the electromagnetic counterparts to BH mergers including the Fermi GBM event after GW150914 are difficult to detect with the current sensitivity. The accretion from the ISM can evaporate the cold neutral dead disc around the BH. A slight accretion before the merger can also blow away the surrounding matter if any. These should be considered as constraints on dead disc models for prompt electromagnetic counterparts of the BHBH merger. Although we do not go into detail in this paper, there are several sites of particle acceleration for a BH jet. First, the BH magnetosphere acts as a particle accelerator like pulsars if a gap
Paper: 2019MNRAS.486.4463Y (fulltext)
could regulate SFRs to some extent (Shi et al. 2011 , 2018 ). 4.2 Theoretical models to explain the relation between σ<sub>gas</sub> and SFR There are two dominant models for the origin of the turbulence: star formation feedback or gravitational instability of the gas (Krumholz Burkhart 2016 ). Both models predict that the velocity dispersions will correlate with their SFRs. The work of Krumholz Burkhart ( 2016 ) gives the details about these two models. In Fig. 8 , we overlaid the gravitational-instability model (left-hand panel) and star-formation-feedback model (right-hand panel). The gravitational-instability model gives SFR as following (Krumholz Burkhart 2016 ): where v c is the rotational velocity, f g is the gas fraction, ε ff = 0.01 is the SFR per free-fall time, φ P = 3 is a factor that accounts for the presence of stars in the disc, and a Coulomb logarithm-like term ln ( r 1 / r 0 ), which measures the radial extent of the star-forming disc ( r 1 = 10, r 0 = 0.1). From equation ( 3 ) we find that a gravitational-instability model has a strong dependence on the gas fraction. The star formation feedback model gives the SFR as following (Krumholz Burkhart 2016 ): where Q g is the Toomre parameters of the gas, φ ≈ 1 and are constants of order unity that parametrize various uncertainties, P * / m * = 3000 km s − 1 is the momentum per unit mass (Faucher-Giguère, Quataert Hopkins 2013 ). We found that the star formation feedback origin for the turbulence predicts that a velocity dispersion rises more steeply with the SFR, and it does not depend on the gas fraction. For the comparisons with our observations, we give the plausible range of f g and Q g following with the works of Krumholz Burkhart ( 2016 ) or Johnson et al. ( 2018 ). We also used the rotation velocity v c = 90–190 , which spans the plausible range for local and high z galaxies. From Fig. 8 (right panel), we find that a star-formation-feedback model provides a rather poor match to the observation, while the gravitational-instability model shows better agreement with the observations. Johnson et al. ( 2018 ) also test two analytic models but found that both provide an adequate description of the data, and need further observations to rule out either model. In the work of Krumholz Burkhart ( 2016 ), they also compared with observations and found that gravity is the ultimate source of ISM turbulence, at least in rapidly star-forming, high-velocity dispersion galaxies for which our test is most effective. Figure 8. The relationship between SFR and gas velocity dispersion. In the left-hand panel, lines show the predictions of the gravitational-instability model (equation 6) for f g = 0.2, 0.5, and 1.0, as indicated in the legend. Lines in the right panel show the prediction of the star formation feedback model (equation 8) for Q g = 0.5, 1.0, and 2.5. The solid lines are for a circular velocity v c = 140 km s −1 (the median rotation velocity of local and high z galaxies), and the
Paper: 2020ApJ...890....2A (fulltext)
regular observations of SN 1987A continued for the purposes of monitoring the evolution of the interaction with the ER and to develop a clear picture of the evolving relationship between the IR emission at these wavelengths and emission at optical and X-ray wavelengths (Dwek et al. 2008 , 2010 ; Arendt et al. 2016 ). This paper reports on the complete set of Spitzer ’ s IRAC observations of SN 1987A and its ER. Section 2 briefly describes the observations and additional mosaicking procedures. The complete mid-IR light curves from Spitzer IRAC and MIPS photometry are presented and modeled with a simple empirical function in Section 3 . The final light curves extend a little over 4 yr (in eight epochs) beyond the results presented by Arendt et al. ( 2016 ). In Section 4 we employ high-resolution mapping and deconvolution that are enabled by the long-term repeated observations of SN1987A at a wide range of position angles. We model the marginally resolved ER to identify the separate evolution of the N, S, E, and W portions of the ER. Section 5 provides further discussion of some of the results, and the findings are summarized in Section 6 . The Appendix describes a strongly variable source that is unassociated with SN 1987A, but was serendipitously located within the field of view of the IRAC observations. 2. Data 2.1. Standard Post-basic Calibrated Data (pBCD) Mosaics and Photometry The initial targeted IRAC observations of SN 1987A employed 12 s frame times using the small-scale Spiral16 dither pattern. Guided by those results, subsequent SN 1987A observations continued the use of 12 s frame times, but switched to the slightly shallower medium-scale Reuleaux12 dither pattern, resulting in 125 s of total exposure (10.4 s of signal integration per 12 s frame time). These observations were repeated at roughly 6 month intervals. The basic calibrated date (BCD) individual frames were automatically mosaicked into pBCD images on 06 pixel scales. The SN brightness was measured from the pBCD mosaics using aperture photometry. Because the SN is not well resolved from nearby stars (Star 2 and Star 3 as designated by Walker Suntzeff 1990 ), the source aperture includes the SN and these stars, and estimates of their flux density (extrapolated from the observations by Walborn et al. 1993 ) are subtracted from the reported result for SN 1987A (Table 1 ). Table 1 SN 1987A Flux Densities Notes. Flux densities are in millijansky. Flux densities of [0.41, 0.26, 0.16, 0.09, 0.01] mJy have been subtracted at [3.6, 4.5, 5.8, 8, 24] μ m to account for the emission of Stars 2 and 3 in the aperture. This correction was not included in Table 1 of Arendt et al. ( 2016 ). a Day number 0 = 1987 Feb 23. b Spitzer astronomical observation request (AOR) number. c Spitzer program ID number. Bold numbers indicate programs specifically targeting SN 1987A. (This table is available in machine-readable form.) A machine-readable version of the table is available. Additional
Paper: 2017MNRAS.464.2462G (fulltext)
a with σ = 0, Ω m = 0.26, and Ω r = 8.4 × 10 −5 . We see that k cr decreases in time and reaches the minimum at a M = [Ω m /2(1 − Ω m )] 1/3 ≃ 0.56. The super-interaction effect appears in a very small region (large k ) in the very early time of the universe. In addition, the larger the ν is, the wider the influences of dark energy to the matter density perturbation will be. Consequently, the size of the super-interaction regime enlarges in the cosmological evolution and finally reaches the scale , 2.8 × 10 −2 , and 2.8 × 10 −1 ( h Mpc −1 ) at a = a M for ν = 10 −2 , 10 −4 , and 10 −6 , respectively. Figure 1. The critical scale factor a cr as a function of the wavenumber k with σ = 0 and ν = 10 −2 (solid line), 10 −4 (dashed line), and 10 −6 (dotted line), where the boundary conditions of Ω m = 0.26 and Ω r = 8.4 × 10 −5 are used. 4 EVOLUTIONS OF MATTER DENSITY PERTURBATION AND POWER SPECTRUM The Hubble radius d H ≡ H −1 increases during the evolution of our universe. More and more density perturbation modes of δ k with the wavenumber k enter the horizon, in which k 2 = a 2 H 2 with a = a k . Explicitly using Ω m = 0.26, Ω r = 8.4 × 10 −5 , and H 0 = 70 km s −1 · Mpc, we find that k = 10 −3 and 0.25 ( h Mpc −1 ) go to the horizon at a k ≃ 0.03 and 1.24 × 10 −5 , respectively. Following a similar procedure to that discussed in the literature (Borges et al. 2008a ), we can recover P ( k ) in the ΛCDM model with the BBKS transfer function within 10 per cent accuracy by taking the initial density perturbation to be the scale invariance when all modes are located at the super-horizon scale, i.e. at a ≪ a k , where n s is the spectral index, given from the inflationary epoch. From the above discussion, it is reasonable to choose ln a = −18 to be the initial time in our work. We note that from equation ( 26 ), the super-interaction effect appears at the scale deep inside the horizon , indicating that the ΛCDM initial condition is applicable to the RVM. In this section, we analyse the evolution of δ m from equation ( 25 ) for two cases: (i) σ = 0 and (ii) σ ≠ 0, and numerically solve δ m ( a ) and P ( k ). (i) σ = 0. At the super-interaction regime ( k ≫ k cr ), the friction in equation ( 25 ) plays the most important role in the evolution history of δ m . Thus, with 0 ν ≪ 1, the growth of δ m is frozen with , and the matter power spectrum is suppressed by the dark energy perturbation. From equations ( 17 ) and ( 18 ), one can derive that and θ m ≃ 0 at k ≪ k cr , showing that δρ Λ = ρ Λ δ Λ 0. When entering the super-interaction scale, the second term on the right-hand side of equation ( 21 ) dominates, and the momentum perturbation of θ m becomes of the order of , implying that the RVM heats up the cold dark matter for a large value of k . As a result, the particles propagate inside the super-interaction divide, and the growth of δ m is frozen. In Fig. 2 (a), we show the evolution of as a function of the scale factor a with k = 0.25 ( h Mpc −1
Paper: 2020MNRAS.494.3937B (fulltext)
modern stellar atmosphere models were compared to the disentangled spectra of each component of HD 54662 AB in order to estimate their physical parameters. A previous quantitative spectroscopic study of the system components was performed by MMR18 using the cmfgen code (Hillier Miller 1998 ). Regarding the projected rotational velocities, using the Fourier Transform method (Simón-Díaz Herrero 2007 ) MMR18 determined 158 and 43 for the genuine 12 primary (A) and secondary (B) components, respectively. These values perfectly agree with our determinations of v sin i (see Table 6 ). However, we differ with the former authors regarding the macroturbulent velocity: while MMR18 attributed to component B a macroturbulence of 35 in order to match the observed line profiles with the synthetic spectra, we estimated a macroturbulence of 40 for component A and fixed this parameter to zero for component B, since we found no reason to modify its value when checking the broadening of the synthetic line profiles. The values computed in this work for the effective temperatures ( T eff ) and surface gravities (log g ) of the components are consistent, within the uncertainties, with the current calibrations for O stars in the Milky Way, including the classical calibration by Martins et al. ( 2005 ), as well as the most recent ones by Simón-Díaz et al. ( 2014 ) and Holgado et al. ( 2018 ). MMR18 also determined T eff , as well as helium abundances ( Y He ) for both stars. While our estimates of log g and Y He are in good agreement with their values, our T eff determinations result in slightly higher values than theirs (Δ T eff 1500 K). We re-analysed both stars using the line broadening parameters estimated by MMR18 . We obtained good agreement for component A, in contrast to component B. Moreover, the broadening of the fastwind synthetic lines using MMR18 ’s v sin i and v mac values cannot reconcile the broadening of the lines observed in our disentangled spectra. Therefore, line broadening only cannot explain the differences in T eff with MMR18 . In Section 3.1 , we mentioned that the presence of residuals in the Balmer lines originated from the disentangling process may affect the surface gravity determination. As explained in Section 5 , for our analysis we used the least affected lines (H γ and H ε) in order to minimize this effect. Regarding the He lines, some profile distortions which could be playing a role are also observed, in particular in the He i λ 4388 Å line, mostly for the primary component. In any case, Fig. 6 shows that the synthetic composite spectrum resulting from the addition of the individual models for the A and B components reproduces well the observed spectrum of the system. We then conclude that potential differences in the shape of the lines of the individual spectra due to different disentangling processes may be the cause of the disagreement between our determinations and those from MMR18 . Note also that, unlike the results presented
Paper: 2018MNRAS.481.4602L (fulltext)
(the second average). This means that the short-term effects in the time-scale of the outer binary’s period are incorporated into the double-averaged Hamiltonian, but the short-term effects in the time-scale of the inner binary’s period are neglected. Thus, the application conditions require that the elements of the test particle are slow variables in the time-scale of the inner binary’s orbital period (i.e. the secular frequencies of the elements need to be much smaller than the mean motion of the inner binary). It is evident that the modified approach broadens the application range compared to the classical double averaging. As for the computational efforts about the classical and modified double-averaged models, we report in Table 1 the time required by the numerical integration of different-order models over a time span of 1.5 × 10 4 yr for the Sun-perturbed Uranus–Caliban system (the initial conditions and time span are the same as the ones of Fig. 6 ). It is observed that, compared to the classical approach, the modified version requires additional computation time. Therefore, when the application conditions of the classical approach are satisfied, the classical averaged model is preferable because in this case the periodic corrections considered in the modified approach become negligible but the computation time increases. Table 1. Computation time for the numerical integration of different-order averaged models about the Sun-perturbed Uranus–Caliban system. The initial conditions and time span are the same as the ones of Fig. 6 . The computation time about the full model over the same period is 9.720 s. In some physical systems, such as Jupiter’s irregular satellite systems, the strong perturbation coming from the disturbing body (the Sun) could make the elements of test particles no longer slow variables in the time-scale of the inner binary’s orbital period. In this case, the short-term effects appearing in the first average could accumulate and produce significant errors to the secular evolution. To overcome this problem, the short-term effects would need to be incorporated also in the first average (over the inner binary’s orbital period). 6 CONCLUSIONS In this work, we developed a modified double-averaging approach, which was used to formulate a more accurate Hamiltonian in comparison to the secular Hamiltonian formulated by the classical double averaging. To test the modified approach, the resulting double-averaged Hamiltonian was applied to a variety of hierarchical triple systems, covering planetary, and stellar scales. In this approach, the disturbing function appearing in the Hamiltonian is averaged twice in order to eliminate the short-term oscillations induced from the orbital motions of inner and outer binaries. When the perturbation coming from the disturbing body is relatively strong, the short-term effects neglected in the classical double averaging may accumulate over long time-scales. This leads to the