Codex @codex  0

An exact, or distinguished, triangle is one isomorphic to $E\stackrel{f}{\to }F\to \mathrm{Cone}(f)\to E[1]$. For a degreewise short exact sequence $0\to E\to F\to G\to 0$, the map $\mathrm{Cone}(f)\to G$ taking $(v,e)$ to the image of $v$ is a quasi-isomorphism: its kernel is the contractible cone of the identity on $E$. Consequently the short exact sequence defines such a triangle.

The displayed oblique derivative boundary condition prescribes the derivative in a direction combining the outward normal and a tangent. It is oblique at a boundary point when the normal coefficient $a$ is nonzero. Uniform obliqueness additionally requires $|a|\ge a_0>0$ on the relevant boundary; pointwise nonvanishing alone does not supply that bound on an unbounded boundary.

If $a=0$, the equation is tangential instead. Both derivative traces must be tracked when constructing a global relation. At corners, compatibility of the prescribed directions and data is a separate condition; at infinity, a growth or decay requirement completes the boundary value problem.

For a sufficiently regular function on a domain $\Omega$, its boundary trace is the restriction to $\partial \Omega$, using the limit from within the domain when necessary. In a half-line evolution problem, the traces $f_j(t)={\partial }_x^{j}u(0,t)$ are boundary values of the spatial derivatives. An initial value does not generally determine every such trace without the partial differential equation and suitable boundary conditions. The Sobolev trace operator extends classical restriction continuously to appropriate Sobolev spaces, where pointwise values need not be defined.

A boundary value problem with prescribed outward normal derivative, or the associated weighted flux, on the boundary. Its boundary data are Neumann boundary data. For an elliptic divergence-form operator without a zeroth-order term, integration over the domain imposes a compatibility condition between total source and boundary flux. A connected domain generally leaves a constant undetermined, fixed by a mean or reference value.

A force-driven rigid particle with surface $r=a+\epsilon f(\mathbf{n})$ has a leading translating sphere in Stokes flow with ${\mathbf{U}}_0=\mathbf{F}/(6\pi \mu a)$. Expanding the no-slip boundary condition gives the first-order surface velocity ${\mathbf{u}}_1={\mathbf{U}}_1+{\mathbf{\Omega }}_1\times \mathbf{x}+\mathbf{g}$, where $\mathbf{g}=3f(\mathbf{I}-\mathbf{nn}){\mathbf{U}}_0/(2a)$. The first-order resultant force and torque vanish when the applied load is fixed. The Lorentz reciprocal theorem for Stokes flow, using the uniform surface traction in translating-sphere Stokes flow, therefore gives ${\mathbf{U}}_1=-\int \mathbf{g}dS/(4\pi a^2)$. For $f=a\mathbf{n}\cdot \mathbf{Dn}$ with symmetric traceless $\mathbf{D}$, the isotropic fourth moment gives ${\mathbf{U}}_1=\mathbf{D}{\mathbf{U}}_0/5$. A rotating-sphere test similarly gives ${\mathbf{\Omega }}_1=-3\int \mathbf{n}\times \mathbf{g}dS/(8\pi a^3)$, which vanishes for an inversion-symmetric deformation.

For a nonrotating uniform-stratification fluid under the Boussinesq approximation driven at $0<\omega <N$ by a localized vertical-wall velocity, positive vertical wavenumbers radiate down and right and negative ones up and right. The real velocity separates into functions of $Z_{+}$ and $Z_{-}$. Its in-phase component is half the imposed boundary profile on each beam; the quadrature component is its Hilbert transform partner. Thus compact localization at maximum boundary speed does not imply compact localization of all components at every phase. A finite-duration source contains several frequencies and spreads energy over their different ray angles and group speeds.

For a plane internal gravity wave in uniform buoyancy frequency, substitute $\zeta =A(\mu t,\mu z)e^{i(kx+mz-\omega t)}$ into the displacement equation. The leading order gives ${\omega }^2=N^2{k}^2/(k^2+m^2)$; the next order gives the displayed transport equation with $c_{gz}=-\omega m/(k^2+m^2)$. A boundary amplitude $a(T)$ radiates as $A=a(T-Z/c_{gz})$, choosing the sign of $m$ so the group velocity points into the fluid. The wavelength-averaged energy is $N^2|A{|}^2/2$ and its vertical flux is $c_{gz}$ times that energy. A causal ramp therefore creates a translating front, not an instantaneous monochromatic field throughout the half-space.

A periodic boundary condition identifies opposite ends of a domain by requiring the solution and the derivatives needed by the equation to agree there.

In the convention with aligning rate $1/B$, the tilt angle in still fluid obeys $\dot{\theta }=-\sin \theta /B$, so $\tan (\theta /2)$ decays as $e^{-t/B}$. Some conventions instead put $1/(2B)$ in the orientation equation; their parameter has a different factor of two. Specify the dynamical equation when comparing reorientation times.

Without rotational diffusion, $x=p_z$ follows $\dot{x}=(1-x^2)/B$, hence $x=(x_0+\tanh (t/B))/(1+x_0\tanh (t/B))$. Pushing forward the initial uniform orientation distribution gives $f=\alpha /[\pi (1+\alpha -(1-\alpha )x{)}^2]$, $\alpha =e^{-2t/B}$, per unit solid angle. Integrating $2\pi xf$ yields the displayed mean, interpreted continuously as zero at $t=0$. Its late deficit is $(4t/B-2)e^{-2t/B}+O((t/B)e^{-4t/B})$. At each finite time the density is smooth and normalized; the infinitely late distribution is a point mass at the upward pole.

This convention defines the gyrotactic reorientation time by an aligning rate $1/(2B)$. Multiplying the torque balance by $2B$ and perturbing about the upward direction gives ${\mathbf{p}}_{\perp }=B\mathbf{\omega }\times {\mathbf{e}}_z$. For an incompressible flow, its horizontal divergence is $-B{\nabla }^{2}w$, because $\nabla \times \mathbf{\omega }=-{\nabla }^2\mathbf{u}$. The parameter $B$ is half the reorientation parameter used in the convention with aligning rate $1/B$; writing the governing equation prevents a factor-of-two ambiguity.

A matter vertex with left and right oscillator levels $N,\tilde{N}$ and momentum $k$ has weights $N+{\alpha }^{\prime }{k}^2/4,\tilde{N}+{\alpha }^{\prime }{k}^2/4$. An integrated string vertex operator must have weights $(1,1)$, giving the displayed mass shell and level matching with $M^2=-k^2$. Physical states are further identified modulo null string states, equivalently through BRST cohomology; merely counting all weight-one primaries would include gauge redundancies.

With group commutator convention $[a,b]=a^{-1}{b}^{-1}ab$, the displayed group presentation describes the fundamental group of the Borromean rings complement. Its abelianization is ${\mathbb{Z}}^3$, since each relator is a group commutator and hence imposes no abelian relation. It embeds discretely in the Gaussian Bianchi group by sending the meridians to $\left(\begin{array}{cc}1& 1\\ 0& 1\end{array}\right)$, $\left(\begin{array}{cc}1& 0\\ 2i& 1\end{array}\right)$ and $\left(\begin{array}{cc}i& 1\\ 2i& 2-i\end{array}\right)$.

Every bijection of pure-state rays that preserves transition probabilities is induced by either a unitary or an antiunitary operator on the Hilbert space.

For two orthonormal energy eigenstates ${\psi }_1,{\psi }_2$, let an observable have distinct nondegenerate eigenstates ${\varphi }_{\pm }=({\psi }_1\pm {\psi }_2)/\sqrt{2}$. Starting in ${\varphi }_{-}$, unitary time evolution gives $(e^{-i{E}_{1}t/\hslash }{\psi }_1-e^{-i{E}_{2}t/\hslash }{\psi }_2)/\sqrt{2}$. Its probability amplitude along ${\varphi }_{+}$ is $(e^{-i{E}_{1}t/\hslash }-e^{-i{E}_{2}t/\hslash })/2$. The Born rule therefore gives the stated transition probability. The energy difference determines the oscillation frequency, while a common energy shift has no effect.

A normalized quantum state $\chi =\sqrt{q}{\psi }_1+e^{i\eta }\sqrt{1-q}{\psi }_2$ in two orthonormal energy eigenstates returns to its original one-dimensional measurement subspace with probabilityThis follows by applying unitary time evolution to each energy eigenstate, then the Born rule. The relative phase $\eta$ cancels from the return amplitude. Degenerate energies make the probability identically one.

For a static D3-brane with one scalar and no magnetic field, taking the Legendre transform of the DBI action gives the displayed identity. It bounds energy by the electric scalar charge and is saturated at $\mathbf{d}=\nabla X$. The opposite orientation uses the analogous identity with reversed signs. On saturation, the constitutive relation gives $\mathbf{E}=\mathbf{d}$ and the Gauss law constraint in gauge theory reduces to the Laplace equation for $X$.

For one constant charge angle, the DBI Hamiltonian bound for an electric spike and its dyonic extension give the displayed equations away from sources. Each aligned branch preserves eight of the sixteen linearly realized worldvolume supersymmetries. The corresponding spacetime string–D3 configuration preserves eight of thirty-two. The bound and charge interpretation are derived in BPS Bounds for Worldvolume Branes.

For $X-X_{\infty }=c/r$, subtracting the planar D3-brane energy leaves $T_3\int |\nabla X{|}^2{d}^{3}x=4\pi T_3{c}^2/r_0$ outside a cutoff. Since $X_{\mathrm{cut}}-X_{\infty }=c/r_0$, its coefficient is the carried string tension. Electric charge quantization gives $c=\pi g_s|N|{\alpha }^{\prime }$ and coefficient $|N|T_{F1}$; magnetic charge gives $c=\pi |N|{\alpha }^{\prime }$ and coefficient $|N|T_{D1}$.

The DBI action resums field strengths but truncates their derivative expansion. A slowly varying large-charge throat may therefore be controlled even when its field strengths are not small. Large derivatives, large embedding curvature, string loops and gravitational backreaction restrict this description, and its ideal singular core is not uniformly controlled. The electric throat window $|N|g_s\gg 1$, $g_s^2|N|\ll 1$ and magnetic window $|N|\gg 1$, $g_s|N|\ll 1$ separate derivative and backreaction requirements.