Tidal Disruption of a Supernova: Mechanisms & Diagnostics

• TDS is a broad term encompassing distinct mechanisms where tidal forces disrupt binary companions, white dwarfs, or supernova ejecta to produce supernova-like transients.
• In the stripped-envelope channel, the natal kick of a newborn neutron star or black hole disrupts a companion star, rapidly forming a disk that drives super-Eddington outflows and luminous optical flares.
• Additional channels involve white dwarf tidal detonations and supernova ejecta capture by supermassive black holes, each yielding unique light curves, spectral features, and transient timescales.

Tidal Disruption of a Supernova (TDS) is a non-uniform term in current astrophysical literature for transients in which tidal forces act on a star, a binary companion, a white dwarf, supernova ejecta, or post-supernova debris in a way that produces or substantially reshapes a supernova-like phenomenon. Recent work applies the label to at least three distinct channels: a newborn neutron star or black hole that tidally disrupts a binary companion after a stripped-envelope supernova and then powers a luminous transient through super-Eddington accretion winds (Tsuna et al., 6 Jan 2025); a white dwarf tidally disrupted by an intermediate-mass black hole, where sufficiently deep encounters can trigger runaway thermonuclear burning and a combined TDE+SN event (Shcherbakov et al., 2012, Vynatheya et al., 27 Jan 2026); and a core-collapse supernova occurring within $\sim100\,R_g$ of a supermassive black hole, whose ejecta are partially captured, self-intersect, circularize, and accrete as a luminous nuclear flare (Lei et al., 14 Sep 2025). A related observational debate concerns AT2018cow, whose late-time UV plateau has been interpreted as favoring an accretion-disk origin over interacting-supernova models and has been discussed in connection with the TDS framework (Inkenhaag et al., 9 Oct 2025).

1. Terminological scope and conceptual variants

The phrase “Tidal Disruption of a Supernova” does not denote a single mechanism. In the stripped-envelope binary model, the supernova occurs first and the tidal disruption follows when the natal kick of the compact remnant drives it into the companion. In white-dwarf encounters with an intermediate-mass black hole, the tidal disruption and the supernova are simultaneous consequences of a single deep encounter. In the nuclear-supermassive-black-hole scenario, the disrupted component is the expanding supernova ejecta rather than the progenitor star itself. This suggests that current usage is umbrella-like rather than taxonomically uniform (Tsuna et al., 6 Jan 2025, Shcherbakov et al., 2012, Lei et al., 14 Sep 2025).

Usage in the literature	Disrupted or captured component	Characteristic outcome
Post-SN binary TDS	Main-sequence companion of a newborn NS/BH	Hydrogen-poor luminous SN or FBOT-like transient
WD–IMBH TDS	White dwarf	Combined TDE+SN, fallback-powered high-energy transient
Nuclear SMBH TDS	Core-collapse SN ejecta	Long-lived luminous nuclear flare or turn-on changing-look AGN

This non-uniformity has observational consequences. The characteristic compact-object mass can range from a neutron star to an intermediate-mass black hole to a supermassive black hole; the disrupted matter can be hydrogen-rich stellar gas, degenerate white-dwarf material, or metal-rich core-collapse ejecta; and the emergent transient can resemble a hydrogen-poor superluminous supernova, a fast blue optical transient, a gamma-ray burst plus supernova, or an ambiguous nuclear transient. A common structural element across the channels is the conversion of gravitationally bound debris into radiation through fallback, circularization, or disk accretion.

2. Post-supernova companion disruption by newborn neutron stars or black holes

In the model of Tsuna and Lu, hydrogen-poor supernovae of Type Ibc arise from massive stars that lost their hydrogen envelopes, typically due to interactions with a binary companion. The proposed TDS channel begins when the natal kick imparted to the neutron star or black hole remnant brings the compact object to a collision with a main-sequence companion, eventually leading to full tidal disruption (Tsuna et al., 6 Jan 2025).

The nominal tidal radius is

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$

For $M_*\sim10\,M_\odot$ and a $1.4\,M_\odot$ neutron star, the summary gives

$$R_t \simeq 0.5\,R_* \simeq {\rm few}\times10^{11}\ {\rm cm}.$$

Numerical simulations cited in that work find that full disruption occurs for pericenter distances $r_p\lesssim R_*$, with the deepest plunges, $r_p\lesssim0.4\,R_t$, disrupted on the first passage. In practice, the disruption criterion is taken to be $r_{\rm cl}\lesssim R_*$, where $r_{\rm cl}$ is the post-supernova closest approach obtained by integrating the kicked orbit.

After disruption, roughly a fraction $f_{\rm bound}\simeq0.6$–$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$0 of the companion’s mass becomes bound debris. The classical fallback rate is written as

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$1

with the most-bound return time $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$2 of order hours to days for stellar-mass tidal disruptions. Tsuna and Lu instead assume that the bound mass rapidly circularizes into a thick disk with

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$3

and characteristic radius

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$4

For $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$5, the effective accretion timescale is

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$6

and the subsequent disk mass-loss history is approximated as

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$7

with $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$8 and $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$9.

The physical significance of this construction is that the accretion flow is not treated as a weak late-time perturbation. It is a promptly assembled, super-Eddington engine whose characteristic timescale is comparable to the optical diffusion time of stripped-envelope ejecta, allowing direct dynamical and radiative coupling to the supernova outflow.

3. Wind interaction, diffusion, light curves, and rates in the stripped-envelope channel

At super-Eddington $M_*\sim10\,M_\odot$0, the disk launches a continuum-driven outflow. For $M_*\sim10\,M_\odot$1, the kinetic luminosity integrated over the launching region is written as

$M_*\sim10\,M_\odot$2

where $M_*\sim10\,M_\odot$3. The characteristic conversion efficiency from disk accretion to wind kinetic power is

$M_*\sim10\,M_\odot$4

comparable to standard super-Eddington disk winds. The wind velocity scales as $M_*\sim10\,M_\odot$5 (Tsuna et al., 6 Jan 2025).

When this fast wind overtakes the slower supernova ejecta, a wind nebula forms. Under the thin-shell approximation, the shock obeys the momentum-balance relation

$M_*\sim10\,M_\odot$6

The photon diffusion time through ejecta of mass $M_*\sim10\,M_\odot$7 and radius $M_*\sim10\,M_\odot$8 is

$M_*\sim10\,M_\odot$9

with $1.4\,M_\odot$0 for a $1.4\,M_\odot$1 profile.

The radiative-transfer treatment is a one-zone ejecta model coupled to wind injection. In the diffusion-injected limit, the peak luminosity and timescale satisfy

$1.4\,M_\odot$2

For $1.4\,M_\odot$3, $1.4\,M_\odot$4, and $1.4\,M_\odot$5, the quoted scalings are $1.4\,M_\odot$6–$1.4\,M_\odot$7 days and $1.4\,M_\odot$8, corresponding to optical peaks from $1.4\,M_\odot$9 to $$R_t \simeq 0.5\,R_* \simeq {\rm few}\times10^{11}\ {\rm cm}.$$0 mag. For more massive ejecta, $$R_t \simeq 0.5\,R_* \simeq {\rm few}\times10^{11}\ {\rm cm}.$$1, the peak shifts to $$R_t \simeq 0.5\,R_* \simeq {\rm few}\times10^{11}\ {\rm cm}.$$2–$$R_t \simeq 0.5\,R_* \simeq {\rm few}\times10^{11}\ {\rm cm}.$$3 days; for low-mass ejecta, $$R_t \simeq 0.5\,R_* \simeq {\rm few}\times10^{11}\ {\rm cm}.$$4, it shifts to $$R_t \simeq 0.5\,R_* \simeq {\rm few}\times10^{11}\ {\rm cm}.$$5–$$R_t \simeq 0.5\,R_* \simeq {\rm few}\times10^{11}\ {\rm cm}.$$6 days.

The same work estimates event rates through $$R_t \simeq 0.5\,R_* \simeq {\rm few}\times10^{11}\ {\rm cm}.$$7 Monte Carlo kick realizations in binaries with pre-supernova separations $$R_t \simeq 0.5\,R_* \simeq {\rm few}\times10^{11}\ {\rm cm}.$$8–$$R_t \simeq 0.5\,R_* \simeq {\rm few}\times10^{11}\ {\rm cm}.$$9, companion masses $r_p\lesssim R_*$0–$r_p\lesssim R_*$1, remnant masses $r_p\lesssim R_*$2 or $r_p\lesssim R_*$3, and kick distributions given by a Maxwellian with $r_p\lesssim R_*$4 for neutron stars or log-uniform $r_p\lesssim R_*$5–$r_p\lesssim R_*$6 for black holes. For realistic $r_p\lesssim R_*$7–$r_p\lesssim R_*$8 and $r_p\lesssim R_*$9–$r_p\lesssim0.4\,R_t$0, the disruption probability is $r_p\lesssim0.4\,R_t$1–$r_p\lesssim0.4\,R_t$2 for neutron-star remnants and somewhat higher if black-hole kicks are large. Folding in that stripped-envelope SNe Ibc are $r_p\lesssim0.4\,R_t$3 of all core-collapse events gives an overall event rate of $r_p\lesssim0.4\,R_t$4–$r_p\lesssim0.4\,R_t$5 of core-collapse supernovae, roughly compatible with the observed rates of fast blue optical transients and hydrogen-poor (super)luminous SNe.

4. White-dwarf tidal disruption and tidal thermonuclear supernovae

In the white-dwarf channel, the relevant tidal radius is

$r_p\lesssim0.4\,R_t$6

or in the more exact form quoted by Shcherbakov et al.,

$r_p\lesssim0.4\,R_t$7

A basic constraint is

$r_p\lesssim0.4\,R_t$8

so only $r_p\lesssim0.4\,R_t$9 can disrupt a white dwarf rather than swallow it whole. Deep encounters with penetration factor $r_{\rm cl}\lesssim R_*$0 can tidally pinch the white dwarf strongly enough to trigger runaway thermonuclear burning; in that formulation, the resulting “tidal supernova” ejects $r_{\rm cl}\lesssim R_*$1 and synthesizes $r_{\rm cl}\lesssim R_*$2 of $r_{\rm cl}\lesssim R_*$3 (Shcherbakov et al., 2012).

For GRB060218/SN2006aj, the fallback time of the most-bound debris is written as

$r_{\rm cl}\lesssim R_*$4

with fallback rate

$r_{\rm cl}\lesssim R_*$5

The observed soft X-ray light curve, which rises to peak at $r_{\rm cl}\lesssim R_*$6 and then decays, is reported to follow $r_{\rm cl}\lesssim R_*$7 extremely well. A fit with $r_{\rm cl}\lesssim R_*$8 and $r_{\rm cl}\lesssim R_*$9–$r_{\rm cl}$0 gives $r_{\rm cl}$1–$r_{\rm cl}$2. The same study derives comparable masses from a jet-photospheric model, $r_{\rm cl}$3–$r_{\rm cl}$4, and from the host-galaxy scaling relation, yielding three independent estimates all near $r_{\rm cl}$5. Its radiative interpretation combines Comptonized blackbody emission from a jet photosphere with self-absorbed synchrotron from the expanding jet front (Shcherbakov et al., 2012).

High-resolution simulations later extended the white-dwarf channel by modeling the tidal disruption of a $r_{\rm cl}$6 C/O white dwarf by a $r_{\rm cl}$7 intermediate-mass black hole using AREPO plus a 55-isotope nuclear reaction network. The impact parameter is defined as

$r_{\rm cl}$8

with simulated values $r_{\rm cl}$9. Wider encounters, $f_{\rm bound}\simeq0.6$0, lead to standard TDEs, while closer encounters produce combined TDE+SN events; the $f_{\rm bound}\simeq0.6$1 fractions of the disrupted white-dwarf material vary from $f_{\rm bound}\simeq0.6$2 at $f_{\rm bound}\simeq0.6$3 to $f_{\rm bound}\simeq0.6$4 at $f_{\rm bound}\simeq0.6$5. The unbound ejecta masses span $f_{\rm bound}\simeq0.6$6–$f_{\rm bound}\simeq0.6$7, the kinetic energies $f_{\rm bound}\simeq0.6$8–$f_{\rm bound}\simeq0.6$9, and the $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$00 masses $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$01–$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$02. Peak radial velocities reach $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$03, and a central cavity develops with little material below a radial velocity of several $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$04 (Vynatheya et al., 27 Jan 2026).

Radiative-transfer calculations using CMFGEN and LONGPOL recover rise times and peak luminosities comparable to supernovae, but with strong asymmetry. In the spherical-equivalent treatment, the rise times to bolometric maximum are $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$05–$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$06 d and $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$07 ranges from $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$08 to $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$09. In 2D, peak $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$10-band magnitudes range from $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$11 to $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$12 mag, the side view is $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$13–$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$14 mag brighter in $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$15 than the plume direction, and polarization is of order $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$16. At nebular epochs, forbidden lines such as [Ca II] $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$17, [Fe II] 4658, and [O I] $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$18 can be skewed and displaced by up to $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$19–$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$20. These calculations define a specific observational subclass: TDE-triggered thermonuclear supernovae with extreme one-sided asymmetry and strong viewing-angle dependence.

5. Capture of supernova ejecta by supermassive black holes in galactic nuclei

A separate TDS scenario considers a core-collapse supernova exploding very close to a supermassive black hole. In the fiducial hydrodynamical setup summarized by Lei et al., the ejecta have

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$21

and are launched at

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$22

from an SMBH of mass

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$23

The surrounding gas is modeled with the self-similar ADAF solution of Narayan and Yi with $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$24. The three-dimensional simulations use Athena++ in a spherical domain

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$25

with a logarithmic $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$26 grid, Newtonian gravity $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$27, and an ideal-gas equation of state with $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$28 (Lei et al., 14 Sep 2025).

Material is gravitationally captured if its velocity falls below the local escape speed,

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$29

The capture cross section is approximated as

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$30

with

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$31

so that the captured mass can be written as $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$32 with $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$33–$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$34. The bound debris returns on eccentric orbits, self-intersects, and circularizes; the circularization timescale $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$35 is defined by the condition

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$36

in the midplane.

The resulting fallback rate is represented as

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$37

where

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$38

and the late-time exponent is measured to be $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$39–$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$40. The circularization timescale satisfies $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$41 with $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$42–$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$43.

In the fiducial run, the net accretion rate at $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$44 reaches

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$45

at a rise time

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$46

For a $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$47 SMBH and $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$48, the Eddington accretion rate is

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$49

so the peak is $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$50. The luminosity,

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$51

reaches

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$52

and the bright phase can last

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$53

This channel is notable because it is not limited by the standard stellar-TDE swallowing problem at high black-hole mass. The summary states that the peak accretion rate of a typical core-collapse-supernova scenario can exceed the Eddington limit for SMBHs with $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$54, while for $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$55 the flow remains sub-Eddington but can still produce flares with $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$56. The predicted light curves fall into two classes: a quasi-steady plateau-like phase when circularization is efficient, and a pure power-law decline when circularization is incomplete. Strong metal lines from O, Si, and Fe are proposed as a distinguishing spectral feature, and the model is offered as an explanation for ambiguous nuclear transients, extreme nuclear transients, and some turn-on changing-look AGNs.

6. Observational diagnostics, discriminants, and interpretive debates

The stripped-envelope post-supernova TDS model was developed partly to address observational properties that are difficult to reproduce with the conventional magnetar spindown mechanism. In that framework, late-time hydrogen lines can arise because the original hydrogen-rich wind, initially unshocked by the supernova ejecta, becomes optically thin after $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$57 d and is then photo-ionized by the central X/UV source. The summary gives

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$58

about $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$59 of the bolometric luminosity, with the line appearing after the ejecta scattering depth drops to $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$60 at $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$61 d for $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$62–$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$63, or earlier, $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$64 d, if $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$65. Grazing encounters can yield multiple partial passages separated by $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$66 d, producing multi-peak bumps on $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$67–$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$68 d timescales. For low-mass ejecta, $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$69, the model also predicts a mildly relativistic outflow with $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$70–$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$71, early mm/radio emission, and the emergence of a central X-ray source after $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$72 weeks (Tsuna et al., 6 Jan 2025).

The white-dwarf tidal-supernova calculations predict a different diagnostic set. Because the ejecta are highly one-sided, the same event can appear faint and red along the plume direction, but $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$73–$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$74 mag brighter in $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$75 from side-on orientations. Spectra in the photospheric phase can switch from narrow, blueshifted absorption-only features to more symmetric P-Cygni profiles or redshifted features depending on viewing angle, while nebular forbidden lines can show velocity-centroid shifts of $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$76–$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$77 together with a central deficit in emission within $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$78. Polarization near maximum of order $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$79 is another stated signature (Vynatheya et al., 27 Jan 2026).

In the nuclear-supermassive-black-hole scenario, the proposed discriminants are long durations, months-to-decades power-law declines, and heavy-element-rich spectra. Since the debris is supernova ejecta rather than a disrupted solar-type star, strong O, Si, and Fe lines are expected to distinguish TDS from ordinary tidal disruption events, particularly in galaxies with overmassive black holes where normal stellar TDEs are suppressed (Lei et al., 14 Sep 2025).

AT2018cow illustrates the present state of model discrimination in the broader TDS-associated literature. New HST/WFC3-UVIS imaging at $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$80 d in F225W, F336W, and F555W, and at $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$81 d in F814W, shows that in the UV the transient faded by $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$82 mag over $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$83 d, with only F336W fading formally by more than $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$84 between $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$85 d and $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$86 d. The late-time optical bands remain essentially constant, consistent with an underlying extended source. The accretion-disk model used for interpretation assumes a low-mass star, $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$87, tidally disrupted by an intermediate-mass black hole with $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$88, rapid circularization into a geometrically thin, optically thick disk, and a fallback rate $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$89. Re-fitting including the 1900 d point gives

$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$90

with the model matching the UV data without renormalization and yielding $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$91. By contrast, interacting-supernova expectations summarized in the same work predict UV magnitudes of $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$92–$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$93 by $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$94 d, whereas the TDE-disk model remains at $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$95–$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$96 mag. At $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$97–$$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$98 d, the predicted divergence becomes F225W $$R_t \simeq \left(\frac{M_{\rm rem}}{M_*}\right)^{1/3}R_*.$$99 mag for the TDE-disk model versus F225W $M_*\sim10\,M_\odot$00 mag for SN–CSM interaction (Inkenhaag et al., 9 Oct 2025).

The principal interpretive issue is therefore not whether TDS-like phenomena exist in some form, but which physical channel is operating in any given transient. Current literature uses the same label for binary post-supernova disruptions, white-dwarf tidal detonations, and supernova-ejecta capture in galactic nuclei. The shared language reflects common tidal-accretion physics, while the observable outputs remain channel-specific.