Sampling consecutive Cardinal cubic B-splines at the centers of their supports gives diagonal values $2/3$ and adjacent values $1/6$, with all other values zero. Finite truncation gives the displayed B-spline collocation matrix, including its endpoint rows without wrapping or adding boundary conditions. Its inverse has alternating signs. The exact inverse norm of midpoint cubic spline collocation is less than three for each finite dimension and approaches three as the dimension grows.

For the unit-spaced cubic cardinal B-spline on $[0,4]$, the four scalar cubic Bernstein basis coefficient rows are $(0,0,0,1/6)$, $(1/6,1/3,2/3,2/3)$, $(2/3,2/3,1/3,1/6)$ and $(1/6,0,0,0)$. Each row applies on its interval with local parameter in $[0,1]$. They result from the compactly supported $C^2$ truncated-cube combination $\frac{1}{6}{\sum }_{j=0}^4(-1{)}^j(\genfrac{}{}{0ex}{}{4}{j})(x-j{)}_{+}^3$. Each graph piece is a cubic Bézier curve with abscissa controls $j,j+1/3,j+2/3,j+1$.

The positive Liouville volume for the base-orientation/positive-fibre orientation is $\mu ={\pi }^{\ast }{\Omega }_a\wedge \omega$. Its fibre integral is $2\pi {\Omega }_a$, so its total volume is $2\pi \mathrm{Area}(M)$. With the stated coframe convention, $\alpha \wedge d\alpha =-\mu$; the signed contact volume and the positive measure should not be confused. The fields $X,H,V$ preserve this volume by their structure equations.

On an oriented Riemannian surface, the vertical vector field generates positive rotation in each fibre of its unit tangent bundle: $Vf(x,v)={{\partial }_{t}f(x,\cos tv+\sin tiv)|}_{t=0}$. Hence $V(v)=iv$ and $V(iv)=-v$. In the canonical frame, $[V,X]=H$, $[V,H]=-X$ and $[X,H]=KV$.

If $h(t)={\int }_0^t{g}_{s}ds$ with deterministic $g\in L^2(0,\infty )$, the finite-time Cameron-Martin theorem densities have second moments $e^{{\int }_0^t{g}_s^{2}ds}$ bounded uniformly in time. Their uniformly integrable martingale limit is the strictly positive displayed density, giving equivalence on the full infinite continuous-path space. Finite-time densities alone do not guarantee such a global density: a nonzero constant drift has infinite energy and its infinite-horizon path law is singular, distinguished by the limiting ratio $B_t/t$.

For a minimally coupled massless scalar field in flat de Sitter spacetime, $a=-1/(H\tau )$ and the Canonically rescaled de Sitter scalar mode is $u_k=a{f}_k$. The equation $u_k^{\prime \prime }+(k^2-2/{\tau }^2)u_k=0$ and the positive-frequency condition $u_k\sim e^{-ik\tau }/\sqrt{2k\mathcal{V}}$ select the displayed mode. Direct differentiation verifies its Wronskian normalization, which ensures the equal-time canonical commutation relations with box-normalized creation and annihilation operators. A conformally coupled massless scalar has a different mode equation.

For Brownian motion started at zero, let $f_n$ equal $|x|$ outside $(-1/n,1/n)$ and $n{x}^2/2+1/(2n)$ inside. Its derivative is the clipped sign function $g_n(x)=\max \{-1,\min \{1,nx\}\}$. The Itô integral $M_t^{(n)}={\int }_0^t{g}_n(B_s)d{B}_s$ converges uniformly on bounded time intervals in $L^2$ to a continuous martingale $M$, by the Itô isometry and Doob L2 maximal inequality: the squared maximal error is at most $8\sqrt{2T/\pi }/n$. The Itô formula gives $\frac{n}{2}{\int }_0^t{1}_{\{|B_s|<1/n\}}ds=f_n(B_t)-f_n(0)-M_t^{(n)}$. Along $n=k^2$, the maximal errors are summable, so First Borel-Cantelli lemma gives uniform almost-sure convergence. The ratio-one interpolation for shrinking-window occupation integrals extends it to every integer $n$. The limit is $|B_t|-M_t=L_t^0$, the local time of a semimartingale in the Tanaka formula convention. The unnormalized factor $n$ gives twice this limit.

The Brownian reflection principle gives $\mathbb{P}({\tau }_x\le t)=\mathbb{P}(|N|\ge x/\sqrt{t})$ for a standard normal random variable $N$. This is the distribution function of $x^2/N^2$. An independent Brownian motion observed at this hitting time has the Cauchy distribution with scale $x$.

For $T_{a,b}=\inf \{s:W_s=a+bs\}$ with $a>0$, exponential tilting of the terminal Brownian value gives $e^{{\theta }^{2}t/2}$ times the barrier-survival Probability for Brownian drift $\theta -b$. Thus the moment equals $e^{{\theta }^{2}t/2}[\Phi ((a+(b-\theta )t)/\sqrt{t})-e^{2a(\theta -b)}\Phi ((-a+(b-\theta )t)/\sqrt{t})]$ for every real $\theta$.

Unit-radius Brownian exit time has finite variance, since independent unit-time increments give an exponential upper tail. The Brownian exit-time skeleton clock after $2^{2n}$ radius-$2^{-n}$ exits has mean $\mathbb{E}{\sigma }_1$ and variance $2^{-2n}\mathrm{Var}({\sigma }_1)$. Chebyshev's inequality and the Borel-Cantelli first lemma yield the displayed pathwise convergence. Independence between different rows indexed by $n$ is unnecessary.

The position is the sum of $k$ independent symmetric signs. Its characteristic function after division by $\sqrt{k}$ is $\cos (u/\sqrt{k}{)}^k\to e^{-u^2/2}$. The Lévy continuity theorem gives the normal distribution limit. Scaling the same skeleton to radius $2^{-n}$ and using Brownian exit-skeleton clock convergence identifies its other limit as $B_{\mathbb{E}{\sigma }_1}$, forcing $\mathbb{E}{\sigma }_1=1$.

For a connected open set $D$, finiteness of the Brownian exit time expectation at one point implies finiteness everywhere. The Brownian exit-time ball averaging identity gives integrability on a ball centred at any finite point. Every point of that ball has a smaller centred ball contained in it; the same identity then gives a finite value there. The set of finite points is open. It is also relatively closed: near a limit point in $D$, choose a fixed ball radius smaller than a quarter of its distance to $D^c$ and centre it at a nearby finite point. The limit point lies in that ball and hence is finite. Connectedness completes the proof.

For a split block of a group algebra with cyclic defect group of a block of order $q=p^n$ and inertial index of a block $e$, its Brauer tree algebra has $e$ simple modules and exceptional multiplicity $(q-1)/e$. Its stable Auslander–Reiten quiver is $\mathbb{Z}{A}_{q-1}/⟨{\tau }^e⟩$, giving $e(q-1)$ nonprojective indecomposables. Adding the $e$ indecomposable projectives gives $eq$ indecomposables in total.

Use a finite tree with cyclically ordered incident edges at each vertex, and multiplicities one except possibly at one vertex. Quiver vertices are tree edges; arrows follow each cyclic successor order, including loops at vertices of valency one. For an edge $i$ with endpoints $v,w$, write $C_{i,v}$ for the cycle around $v$. Impose $C_{i,v}^{m_v}=C_{i,w}^{m_w}$, $C_{i,v}^{m_v}{\alpha }_{i,v}=0$, and zero for any path that switches from one vertex-cycle to the other. The quotient of this path algebra is the basic Brauer tree algebra; redundant valency-one loops can be eliminated. Its indecomposable projective at $i$ has simple top and socle labelled by $i$, with two uniserial radical branches of lengths $m_v\mathrm{val}(v)$ and $m_w\mathrm{val}(w)$ meeting in that socle.

Let $i$ be an edge at $v$. Kill the other radical branch of its indecomposable projective module. The resulting uniserial module has length $m_v\mathrm{val}(v)$ and its factors follow the cyclic edge order beginning at $i$. Quotienting its radical powers gives all shorter such modules. The edge succession uniquely fixes the path, and its nonzero arrow maps can be rescaled to one, giving uniqueness up to isomorphism. At an exceptional leaf this reduces to a nilpotent single Jordan chain; the other leaf loop, when present, is a socle path and vanishes in these shorter quotients.

Over a splitting field for finite group representations, blocks of a group algebra with defect group of a block $D$ correspond to $N_G(D)$-orbits of blocks $b_D$ of $k(D{C}_G(D))$ with defect $D$ whose stabilizer quotient $N_G(D,b_D)/(D{C}_G(D))$ has order coprime to $p$. Equivalently one uses the defect-zero block of $k(D{C}_G(D)/D)$. This quotient is the inertial quotient of a block.

In the compactified Riemann surface of $w^r=z^n-1$, a positive circuit enclosing the $n$ simple finite zeros sends the sheet label $j$ to $j+n$ modulo $r$. This permutation has $d=\gcd (n,r)$ cycles of length $r/d$. Each cycle fills in one point over infinity, with ramification index $r/d$. Hence infinity is unramified exactly when $r$ divides $n$. Each finite zero has ramification index $r$, so the base sphere has exactly $n$ branch values in that case and $n+1$ otherwise.

Vernier's Boyle's Law: Pressure-Volume Relationship in Gases, Experiment 6 of Chemistry with Vernier, specifies a graduated $20\mathrm{mL}$ gas syringe, a pressure sensor and LabQuest recording. Preparation sets the disconnected syringe to $10.0\mathrm{mL}$, then attaches it to the sensor valve, trapping air. At several plunger positions, the displayed pressure is allowed to stabilize before the paired volume is entered. The analysis tests whether $pV$ remains approximately constant. It adds $0.8\mathrm{mL}$ to each syringe reading for the sensor's internal volume; this correction belongs to that apparatus. The worksheet supplies a blank results table, not measured values.^[ref]

The manufacturer's Gas Pressure Sensor User Manual identifies a membrane sensor referenced to a vacuum, hence an absolute pressure readout. Its calibration section uses the actual local atmospheric pressure for an ambient reference check, rather than a weather report corrected to sea level. Checking this baseline before sealing separates an offset error from compression physics.^[ref]

Additional controls follow from the model. Keep the same sealed amount of substance, connections and ambient temperature; persistent pressure drift at held volume after thermal equilibrium suggests a leak or another uncontrolled change. Small steps and waiting allow compression heat to leave the gas; avoiding hand warming reduces another heat input. Compare settled readings when approaching a position from either direction. Plunger friction affects the applied force and position reproducibility, so record actual pressure and actual plunger position rather than predicting pressure from the load. These are proposed controls, not reported experimental outcomes.

Nearly constant $pV$ across equilibrated states supports the fixed-temperature ideal gas approximation over the tested range. Stabilized pressure alone does not demonstrate uniform or constant temperature; the isothermal gas syringe experiment adds a temperature readout to distinguish a transient from its cooled endpoint. Finite sensor accuracy and uncertain total volume limit the test, and agreement does not establish ideal behavior at every pressure. Repeated return readings help distinguish reproducibility from a one-direction trend.

A linearized plot of Boyle's law uses $x=1/V$ as its horizontal coordinate. Then absolute pressure satisfies $p=Cx$: the slope is $C=nRT$, and the ideal intercept is zero. In the figure, $(V,p)=(100,100)$ and $(50,200)$ in $\mathrm{c}{\mathrm{m}}^3$ and $\mathrm{kPa}$ become $(x,p)=(0.010,100)$ and $(0.020,200)$ in $\mathrm{c}{\mathrm{m}}^{-3}$ and $\mathrm{kPa}$. Both give slope $10000\mathrm{kPa}\mathrm{c}{\mathrm{m}}^3$. The Vernier experiment includes this reciprocal-volume graph as its extension.^[ref]

At constant atmospheric pressure, gauge pressure instead obeys $p_g=Cx-p_{\mathrm{atm}}$, with the same slope and a negative intercept. A nonzero fitted intercept in an intended absolute pressure plot can indicate sensor bias or an incorrect volume correction, but is not by itself a unique diagnosis. The origin is a mathematical extrapolation to infinite volume and zero absolute pressure, rather than an additional measured state.

With gauge pressure $p_g$ referenced to atmospheric pressure $p_{\mathrm{atm}}$, absolute pressure is $p=p_g+p_{\mathrm{atm}}$. Thus Boyle's law becomesEven a constant atmospheric pressure cannot be dropped from these sums. For the calculated $100\mathrm{kPa}$ to $200\mathrm{kPa}$ example, take $p_{\mathrm{atm}}=100\mathrm{kPa}$. The initial and final gauge pressures are $0$ and $100\mathrm{kPa}$; their products with the respective volume are $0$ and $5000\mathrm{kPa}\mathrm{c}{\mathrm{m}}^3$, which are unequal. Zero gauge pressure means balance with the surroundings, not absence of gas or absence of wall momentum transfer.

For a frictionless piston of surface area $A$ with its outer face exposed to atmospheric pressure, the additional inward force required for mechanical equilibrium is $F=(p-p_{\mathrm{atm}})A=p_{g}A$. This is why a mechanical load can reveal a pressure difference even though the equation of state requires absolute pressure. With piston friction, the load alone no longer determines $p_{g}A$; a direct pressure sensor avoids that particular inference.