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Applying the product of the other eigenvalue factors to a linear relation isolates each coefficient. Distinct eigenvalues therefore give linearly independent eigenvectors. A complete eigenvector basis diagonalizes the associated linear system of ordinary differential equations.

An adjoint eigenfunction is a nonzero eigenfunction of the adjoint operator on its adjoint domain. For an inner product conjugate-linear in its first argument, $L^{\ast }g=\overline{\lambda }g$ implies $⟨g,(L-\lambda )f⟩=0$ for every $f$ in the domain of $L$. Consequently a necessary solvability condition for $(L-\lambda )f=h$ is $⟨g,h⟩=0$. If $L-\lambda I$ is a Fredholm operator, orthogonality to every element of its adjoint kernel is also sufficient. Boundary conditions and weights in a coupled system enter the adjoint domain and must be derived by integration by parts; copying the original eigenfunction without checking them can give a wrong projection.

For a vertex set of size $m$ in a hypercube graph, split one coordinate into sections of sizes $a,b$. Induction bounds the internal edges within the sections by $(a{\log }_{2}a+b{\log }_{2}b)/2$; at most $\min (a,b)$ edges cross between them. Put $t=\min (a,b)/(a+b)$. The chord bound for the binary entropy function gives $H_2(t)\ge 2t$, hence $a{\log }_{2}a+b{\log }_{2}b+2\min (a,b)\le m{\log }_{2}m$. Thus at most $m{\log }_{2}m/2$ internal edges are present, and the edge boundary is at least $m(n-{\log }_{2}m)$. Coordinate subcubes attain equality when $m$ is a power of two.

The clique conflict graph has one vertex for each $r$-clique of a fixed graph, and joins two when they share an edge. Its independent sets correspond exactly to edge-disjoint clique packings. The Caro-Wei bound converts clique-count and overlap estimates into a packing lower bound.

Twice-applied spin-raising or lowering annihilates the ordinary spherical-harmonic multipoles $\ell =0,1$. Polarization potentials therefore have a four-dimensional monopole/dipole ambiguity on the sphere. For an axisymmetric mode it is $A+B\cos \theta$. Setting these components to zero defines the observable $\ell \ge 2$ potentials without changing $Q$ or $U$.

For a scalar photon quadrupole emitted at distance ${\chi }_{\ast }$, the meridian-basis source is $Q\pm iU\propto {\Theta }_2{e}^{-ik{\chi }_{\ast }\mu }(1-{\mu }^2)$. Integrating its scalar-potential equation and removing the low-multipole kernel of polarization potentials yields $P_E\propto {\sum }_{\ell \ge 2}(-i{)}^{\ell }(2\ell +1){\Theta }_2{j}_{\ell }(k{\chi }_{\ast })P_{\ell }(\mu )/(k{\chi }_{\ast }{)}^2$, with no scalar $P_B$. The apparent $k^{-2}$ divergence lies in discarded low multipoles; $j_{\ell }(x)=O(x^{\ell })$ makes the physical small-$x$ limit regular.

A mean-field dynamo evolves a slowly varying spatially averaged magnetic field. Correlations of fluctuating velocity and magnetic field supply a mean-field electromotive force. Its local linear response includes the alpha effect; gradients of the mean field can contribute additional transport terms. The distinction between a uniform test field used to measure a response coefficient and a spatially varying mean field used to obtain growth is essential.

For an isolated bounded conductor of uniform magnetic diffusivity $\eta >0$, contained in a sphere of radius $R$ and matched to a decaying potential field in an insulating exterior, dynamo action requires maximum stretching rate $S\ge \eta {\pi }^2/R^2$. Here $S$ bounds the largest eigenvalue of the rate-of-strain tensor throughout the flow and time. Use fluid boundary conditions eliminating the stretching boundary term, for example a no-slip boundary condition, and no imposed energy input. The proof combines the magnetic free-decay spectral bound with the magnetic energy equation. This is a necessary condition, not a sufficiency criterion; changing magnetic boundary conditions changes the spectral constant.

For the dynamic boundary condition for the heat equation with $\alpha =2a$, $\beta =a^2$, the boundary resolvent denominator is $(\sqrt{p}-a{)}^2$. With the principal square root, its inverse kernel isEquivalently, $J_a={\int }_0^{\infty }r{e}^{ar}P(x+r,t)dr$ using the Heat Poisson kernel. This representation fixes the repeated-root sign and remains valid for either sign of $a$.

A dyadic interval is one cell in a dyadic partition. Its children are the left and right half intervals. Disjoint dyadic intervals at a fixed level produce orthogonal indicator functions and independent Brownian motion increments.

The replacement $s=2\beta (\theta -\sin \theta )\mapsto 2\beta \theta$ neglects a bounded orbital ripple in accumulated lag. Substituting the resulting angle into the instantaneous periodic drift gives a convenient ${\csc }^2(s/(4\beta ))$ profile, but is not a uniform small-$\beta$ approximation. The consistent first-order profile retains the implicit angle-to-lag map; near release it gives $n(s)\propto s^{-2/3}$.

For fixed ${\omega }^2={\Omega }^2-2\pi G{\sigma }_0|k|+c^2{k}^2>0$, the two oscillatory roots of the dust gravitational dispersion relation with gas drag acquire negative real parts at first order in $ϵ$. Their frequencies have no first-order shift. The expansion is not uniform near ${\omega }^2=0$, where the oscillation and damping scales can become comparable.

The mode originating at the neutral root grows for $0<{\omega }^2<{\Omega }^2$. At small nonzero $|k|$ this gives $s_{\sec }\sim 2\pi ϵG{\sigma }_0|k|/\Omega$. Gas drag transfers angular momentum to the fixed gas flow and permits secular gravitational instability even for a dust Toomre parameter above unity. If ${\omega }^2<0$, the small-root continuation is damped, while a separate dynamical root grows. The formula is singular at ${\omega }^2=0$.

For any nonzero positive drag, the cubic dust gravitational dispersion relation with gas drag is negative at $s=0$ if ${\omega }^2<{\Omega }^2$ and positive at sufficiently large positive $s$. The intermediate value theorem therefore guarantees a growing real root. This existence argument is independent of the weak-drag expansion and includes the dynamically unstable ${\omega }^2<0$ regime.

Where $C_{KK}>0$, rearranging the Dupire equation gives $\sigma (T,K{)}^2=2(C_T+rK{C}_K)/(K^2{C}_{KK})$. The strike curvature represents discounted density, while the adjusted maturity derivative gives the local diffusion contribution.

For the planar heat equation with point injection of strength $p(s)$, zero initial data gives $\Psi (t,\mathbf{r})={\int }_0^{t}p(s)e^{-r^2/[4D(t-s)]}/[4\pi D(t-s)]ds$. If injection is $p_0$ until $t_0$ and zero later, then for $t>t_0$ this equals $p_0[\mathrm{Ei}(-b/(t-t_0))-\mathrm{Ei}(-b/t)]/(4\pi D)$, where $b=r^2/(4D)$. The exponential integral expansion gives $p_0\ln [t/(t-t_0)]/(4\pi D)$ when $b/(t-t_0)\ll 1$.

In a dry galaxy merger, zero orbital energy at infinite separation and retention of all mass and energy give $E_F=E_I+E_A$. If the progenitors and remnant satisfy the virial theorem, their gravitational radii obey $R_F/R_I=(1+\eta {)}^2/(1+ϵ\eta )$, where $\eta =M_A/M_I$ and $ϵ=⟨v_A^2⟩/⟨v_I^2⟩$. The latter average is weighted by satellite mass. Nonzero orbital energy, escaping stars, or dissipative losses change the energy accounting.

For a down-set $D$ in the Boolean lattice, every member $A$ has all its $|A|$ immediate lower neighbors in $D$. Counting an internal edge at its upper endpoint gives $e(D)={\sum }_{A\in D}|A|$. The degree identity for the edge boundary of the hypercube graph then gives the displayed formula.

Two degree-raising differentials square to zero and anticommute. If an initial convention uses commuting differentials, introduce a sign in one of them before forming the total cochain complex.

For a nonnegative martingale, $T_R=\inf \{n:M_n\ge R\}$ satisfies $\mathbb{P}(T_R<\infty )\le \mathbb{E}{M}_0/R$. Stop at $T_R\wedge n$ and use nonnegativity before passing to the increasing event. A limit of thresholds from below also gives the bound for $\{{\sup }_n{M}_n\ge R\}$ even when the supremum is not attained.